content
stringlengths 15
7.12M
| publicationDate
stringlengths 10
10
| abstract
stringlengths 41
2.5k
| title
stringlengths 14
230
| id
stringlengths 11
12
|
---|---|---|---|---|
1 Electron -Phonon Scattering governs both Ultrafast and Precessional Magnetization Dynamics in Co -Fe Alloys Ramya Mohan1, Victor H. Ortiz2, Luat Vuong2, Sinisa Coh2, Richard B. Wilson1,2* 1Materials Science & Engineering Program, University of California, Riverside 2Department of Mechanical Engineering, University of California, Riverside * Corresponding Author : rwilson@ucr.edu Abstract Recent investigations have advanced the understanding of how structure -property relationships in ferromagnetic metal alloys affect the magnetization dynamics on nanosecond time -scales . A similar understanding for magnetization dynamics on femto - to pico -second time -scales does not yet exist. To address this, we perform time -resolved magneto optic Kerr effect (TRMOKE) measurements of magnetization dynamics in Co -Fe alloys on femto - to nano-second regimes . We show that Co -Fe compositions that exhibit low Gilbert damping parameters also feature prolonged ultrafast demagnetization upon photoexcitation. We analyze our experimental TR - MOKE data with the three-temperature -model (3TM) and the Landau -Lifshitz -Gilbert equation . These analyses reveal a strong compositional dependence of the dynamics across all time -scales on the strength of electron -phonon interactions. Our findings are beneficial to the spintronics and magnonics community , and wil l aid in the quest for energy -efficient magnetic storage applications. Introduction Laser excitation of a magnetic metal causes energy to cascade from photoexcited electrons into spin and vibrational degrees of freedom1–3. In ferromagnetic 3d transiti on metals such as Fe, Co, and Ni, the rapid increase in thermal energy stored by spin degrees of freedom causes femtosecond quenching of the magnetization2,3, followed by a partial recover over the next few picoseconds . Subsequently, on nanosecond time-scales , a temperature induced change in equilibrium properties causes oscillatory precessions of the magnetic moment . Both ultrafast and precessional magnetization dynamics involve energy exchange between magnetic and vibrational degrees of freedom . The energy exchange is mediated by quasi -particle interactions . The strength of quasi -particle interactions in a ferromagnet depends on e lectronic band structure4,5. In 3d ferromagnetic alloys, the electronic energy bands near the Fermi -level vary strong ly with composition6. Several recent investigations of nanosecond precessional dynamics in ferromagnetic alloy s have explored the relationship between electronic band structure, quasi -particle interactions , and magnetic damping6–8. Schoen et al. report a n intrinsic 2 damping parameter less than 10-3 for Co 0.25Fe0.756, which is unusually low for a metal . They conclude that the low damping in Co0.25Fe0.75 is a result of a minimization in the density of states at the Fermi -level, which decreases the rate of electron -phonon scattering . Researchers have not yet reached a unified understanding of how quasi -particle interactions govern the magnetization dynamics in the femtosecond regime2,9–15. Some studies have hypothesized that spin-flips caused by electron -phonon interactions are key drivers of femtose cond magnetization dynamics9,11. Other experimental and theoretical studies have explored the importance of electron -magnon interactions12–15. Encouraged by the recent advances in the materials science of nanosecond precessional dynamics6–8, we study the compositional dependence of ultrafast magnetization dynamics in Co -Fe alloys. Our study’s goal is to understand the relationship between electronic band structure , quasi -particle interactions, and femto -magnetism properties of ferromagnetic metal alloys. We perform time -resolved magneto optic Kerr effect (TR-MOKE) measurements to characterize the magnetization dynamics of thin CoxFe1-x alloy films (capped and seeded with Ta/Cu layers on a sapphire substrate) on femto - to nanosecond time-scale s. See Methods for details on sample geometry. We observe that the ultrafast magnetization dynamics are a strong function of Co - concentration , see Figure . 1a. The ultrafast dynamics of Co xFe1-x differ most significantly from those of Co and Fe at a composition of x = 0.25. We also analyze the time -resolved macroscopic precessional dynamics and report the effective damping parameter of our samples , see Figure 2a. After linewidth analyses, f or CoxFe1-x, we observe that the Gilbert damping parameter varies from 3.6 ×10−3 to 5.6 ×10−3 for compositions between x = 0 and 1, with a minimum value of 1.5 ×10−3 at x = 0.25 , in good agreement with previously reported results , see Figure 3b . To determine the strength and composition dependence of electron -magnon and electron -phonon quasi -particle interactions , we analyze our ultrafast magnetization dynamics data with a three- temperature -model (3TM)2,16. Our results reveal a strong composition al dependence of the electron -phonon energy transfer coefficient, 𝑔𝑒𝑝, suggesting that the variation in the ultrafast dynamics in Co xFe1-x alloys occurs primarily due to electron -phonon scattering. We draw this conclusion because t he value of 𝑔𝑒𝑝 depends on the rate of phonon emission by hot electrons 17. Electron -phonon scattering is also predicted to govern the dampin g of nanosecond precessional dynamics 6,18,19. Therefore, o ur results demonstrate that the same microscopic electron -phonon interactions responsible for Gilbert damping also play a dominant role in femto -magnetism properties of ferromagnetic alloys. Results Ultrafast Magnetization Dynamics We plot the normalized u ltrafast magnetization dynamics response , ∆M(t), for Co, Fe, and Co0.25Fe0.75 as a function of time delay in Figure . 1a. Data for the rest of the Co -Fe compositions are plotted in Supplementary Figure 1. All our measurements were performed with an incident 3 laser fluence less than ~15 J/m2. This is a sufficiently small fluence for the dynamics in our experiments to follow a linear regime. In other words, decreasing the incident f luence by a factor of two decreases the optical signal by a factor of two, but does not change the time-dependence of the signal . We use a polar TR -MOKE configuration t o measure the ultrafast magnetization dynamics at femtosecond time delays. A schemati c of our experimental setup is shown in Supplementary Figure 2a. We apply a n external 2.2 Tesla (T) field perpendicular to the plane of the sample using an electromagnet (GMW 3480). This external field is strong enough to effectively overcome the in-plane shape anisotropy of the Co -Fe alloys and saturate the moment in the out - of-plane direction. Since the equilibrium orientation of the moment is in the out -of-plane direction , both, before and after laser irradiation, this geometry allows us to quan tify the femtosecond demagnetization response of the Co -Fe alloys , without the presence of macroscopic precessional dynamics , see schematic in Fig ure 1b. Upon excitation with the pump pulse, the magnetic moment decreases on a sub -picosecond time- scale due to the flow of energy from electrons to magnons2,3,16,20,21. Then, on picosecond time- scale s, the magnetization partially recovers as energy is transferr ed to the lattice and temperature gradients across the film thickness relax. After a few picoseconds, the magnetic film reaches a new equilibrium at an elevated temperature. Ultrafast dynamics with sub -picosecond demagnetization followed by picosecond re-magnetization are commonly categorized as “type I” dynamics , and are characteristic of 3d ferromagnetic metals such as Fe, Co, and Ni9. To elucidate how the de - and re -magnetization dynamics change with composition, we define two data descriptors : τD and R. We define the demagnetization time , τD, as the delay time where d∆M(t)/dt reaches its maximum value. We define R as the ratio of the maximum of 𝛥𝑀(𝑡) to 𝛥𝑀(𝑡≈10ps). We plot τD and R as a function of composition in Figure 3a. τD varies weakly with composition and has a minimum value of 40 fs at x = 0.25. In contrast , we observe that R varies strongly with composition and is a maximum of 4 at x = 0.25. Nanosecond Precessional Dynamics We show measurements of the macroscopic precessional dynamics of Fe, Co, and Co 0.25Fe0.75 in Figure 2a. Data for the other Co -Fe compositions are plotted in Supplementary Figure 3. We use a polar TR -MOKE experimental setup, with an obliquely angled external magnetic field, to measure the macroscopic precessional dynamics of our samples. A schematic of our experimental setup is shown in Supplementary Figure 2b. Tilting the electromagnet to an angle of 11° , with respect to the plane of the sample, allows us to apply a canted external magnetic field so that the magnetic moment has an out -of-plane component. The equilibrium orientation of the moment depends on the balancing between the applied external field and the thin -film shape anisotropy field. The shape anisotropy field in the z -direction is proportional to the out-of-plane component of the magnetic moment. Upon heating, the total magnetic moment decreases . This decrease results in an ultrafast change to the out-of-plane anisotropy field and equilibrium 4 orientation . As a result, t he magnetic moment will precess to a new equilibrium orientation , see schematic in Figure 2b. Our polar TR -MOKE setup detects changes in the out -of-plane moment , so we can sensitively measure the frequency and amplitude of the precessional dynamics. We collect between 6 and 12 TR-MOKE scans of precessional dynamics for each sample . Each of these scans is co llected with a different applied external magnetic field , ranging from 0. 2 T to 2.2 T. The TR -MOKE signals include precessional dynamics in addition with a background related to temperature -induced demagnetization. To analyze the precessional dynamics, we subtract the background with a biexponential decay function . We fit the resulting dataset with a damped harmonic function, V(t)=Asin(ωt+∅)exp (−t/τ). Our fits yield unique values of A (amplitude), ∅ (the initial phase of the oscillation), T (period), and τ (the exponential decay time of the precession). Using these values, we determine the effective dimensionless damping parameter , αeff = ω.τ-1. The resonance frequency is a function of applied external magnetic field and magnetic moment, 𝜔=γ √Heff(Heff+μ0Ms). Here, ɣ is the gyromagnetic ratio, μ0 is the vacuum permeability, Heff is the out -of-plane component of the external magnetic field as measured by a Hall probe , and Ms is the saturation magnetization of the sample . We derive the magnetic moment of the sample by treating Ms as a fit parameter . We also perform VSM measurements of the moment of some of the samples and find that the magnetic moment obtained is in good agreement with the value that we derive by fitting our precessional dynamics data . See Supplementary Figure 4 for more details . The effective damping parameter α eff that we deduce from our precessional dynamics measurements includes effects from damping and inhomogeneous broadening. The effect of inhomogeneous broadening is independent of the applied field at high frequencies22. To obtain the Gilbert damping parameter intrinsic to the sample geometry (not intrinsic to the material) , we plot the effective linewidth, αeff∙f, as a function of frequency, and linearly fit to the equation , αeff∙f=α∙f+∆H, where ∆H is the inhomogeneous broadening component and α is the Gilbert damping parameter . Further details can be found in Supplementary Figure 5. In contra st to prior investigations that performed FMR measurements in the frequency range from 16 -18 GHz8 and 40 GHz6, our TR -MOKE experimental setup allows us to study dynamics at frequencies as large as 90 GHz. At such high frequency, we can be confident that our measured Gilbert damping parameter is dominate d by the intrinsic linewidth over inhomogeneous broadening effects. The Gilbert damping parameter we observe of α = 1.5 ×10−3 for Co 0.25Fe0.75 is amongst the lowest ever reported for a ferromagnetic metal. Schoen et al. report α=2.1 ×10−3 for Co0.25Fe0.75. After accounting for radiative and spin -pumping contributions, they estimate an intrinsic damping parameter for Co0.25Fe0.75 to be αint=5 ×10−4 . Lee et al. 8 performed FMR measurements of Co0.25Fe0.75 epitaxial films and report α=1.4 ×10−3. Wei et al. report α=5 1.5 ×10−3 for Fe 0.75Al0.25 films 7. We note that our measured damping parameter likely includes significant contributions from spin -pumping into the adjoining Ta /Cu layers, but we did not experimentally examine the effect s of spin -pumping in our samples. Analysis and Discussion The c omparison of 𝑅 and 𝛼 in Figure 3a and Figure 3b reveals that the two quantities depend on composition in a similar manner. R is at a maximum and 𝛼 is at a minimum at x = 0.25 . Fe and CoxFe1-x alloys with x ≥ 0.5 have small R and high 𝛼. Alternatively, C oxFe1-x alloys with 0.1< x < 0.5 have both high 𝑅 and low 𝛼. To confirm this correlation , we performed a hierarchical cluster analysis of the raw data at both femtosecond and nanosecond time-scale s. The clustering algorithm divides the Co -Fe alloys into groups based on similarit ies in the dynamics data . The clustering results as a function of composition are nearly identical when based on the femto - /pico -second time -scale data vs. the nanosecond time -scale data. We include further details on the clustering analysis in Supplementary Note 1 and Supplementary Figure 6 . We now explain the correlation between ultrafast and precessional dynamics by considering how electronic scattering processes depend on composition. Similar to prior studies of damping in Co-Fe alloys6,7,23, our results for 𝛼 vs. x are in good agreement with the “breathing Fermi surface ” model for damping24. In this model , spin -orbit coupling causes the Fermi -level to shift with the precessi ons of the magnetic moment25. A shift in the equilibrium Fermi -level leads to a nonequilibrium electron population . As the Fermi -level repopulates, i ntra-band electron -phonon scattering transfers energy to the lattice . The “breathing Fermi surface” model predicts that the damping parameter is directly proportional to 𝐷(𝜀𝑓), because more electronic states near 𝜀𝑓 leads to higher rates of electron -phonon scattering . We observe that the 𝛼 value for Co0.25Fe0.75 is ~2.5x lower th an 𝛼 for Fe. Density functional theory predicts a ~2x difference in 𝐷(𝜀𝑓) for Co0.25Fe0.75 vs. Fe, see Supplementary Note 2 or Ref.6. Therefore, like prior studies of Co -Fe alloys6,7,23, we conclude that intra -band electron -phono n scattering governs precessional damping. To better understand how composition affects electron -magnon and electron -phonon energy transfer mechanisms , we analyze our 𝛥𝑀(𝑡) data with a phenomenological three temperature model (3TM) , see Figure 4. The 3TM describes how heat flows between electrons, phonons, and magnons after laser excitation of the Co-Fe sample . (See Methods for additional details. ) The 3TM predicts that τD depends on two groupings of model parameters: 𝜏𝑒𝑚≈𝐶𝑚/𝑔𝑒𝑚 and 𝜏𝑒𝑝≈ 𝐶𝑒/𝑔𝑒𝑝. Here 𝐶𝑚 and 𝐶𝑒 are the magnon and electron heat-capacity per unit volume, and 𝑔𝑒𝑚 and 𝑔𝑒𝑝 are the energy transfer coefficients from electrons to magnons an d phonons, respectively. We estimate v alues for 𝐶𝑒 vs. composition using the Sommerfeld model together with the electronic density of states vs. composition reported in Ref.6. The 3TM also predicts that the parameter R is determined by the following grouping of parameters: 𝑅= 𝐶𝑝𝑔𝑒𝑚/𝐶𝑚𝑔𝑒𝑝 16, where 𝐶𝑝 is the phonon heat -capacity per unit volume . We assume that the value of 𝐶𝑝 is 3.75 6 MJ m-3 K-1 for Co, Fe and Co -Fe alloys. With these estimates for 𝐶𝑒 and 𝐶𝑝, and other relevant model parameters, summarized in Supplementary Table 1, we can deduce unique values for 𝐶𝑚/𝑔𝑒𝑚 and 𝐶𝑝/𝑔𝑒𝑝 as a function of composition from our TR-MOKE data, see Figure 4b. Based on our 3TM analysis, we conclude that the strong composition dependence of R is due to the composition dependence of 𝑔𝑒𝑝. Boltzmann rate -equation modelling of the nonequilibrium electron dynamics after photoexcitation predicts that the electron -phonon energy -trans fer coefficient is 𝑔𝑒𝑝=[𝜋ℏ𝑘𝐵𝐷(𝜀𝐹)]𝜆⟨𝜔2⟩ 5. Here, 𝜆⟨𝜔2⟩ is the second frequency moment of the Eliashberg function and is a measure of the strength of electron -phonon interactions . Most of the composition al dependence we observe in 𝑔𝑒𝑝 is explained by the composition al dependence of 𝐷(𝜀𝑓). To show this, we include a prediction for 𝑔𝑒𝑝 in Figure 4b. Our prediction uses the 𝐷(𝜀𝑓) vs. x reported in6 and treats 𝜆⟨𝜔2⟩ as a composition independent fit parameter . We find 𝜆⟨𝜔2⟩=260 meV2 provides an excellent fit to our data . The best-fit value for 𝜆⟨𝜔2⟩ is in good agreement with 𝜆⟨𝜔2⟩≈𝜆𝑅Θ𝐷22⁄=280 meV2. Here, 𝜆𝑅 is derived from electrical resistivity data for Fe 26, and Θ𝐷=470𝐾 is the Debye temperature of Fe. Before beginning our experimental study, we hypothesized that the energy transfer coefficient between electrons and magnons, emg , would be correlated with the phase -space for electron - magnon scattering . We expected the phase -space for electron -magnon scattering to be a strong function of band -structure near the Fermi -level 12–15. We also expected the phase -space to be minimized at a composition of x = 0.25, because of the minimum in the density of states at the fermi -level. To explore how the phase -space for electron -magnon scattering depends on composition, we performed density functional theory calculations for the electronic band structure with x = 0 and x = 0.25, see Supplementa ry Note 2. Our DFT calculations suggest that the phase -space for electron -magnon scattering is an order of magnitude higher for x = 0 vs. 0.25. However, we do not see evidence that this large theoretical difference in electron -magnon scattering phase -space affects ultrafast dynamics . The time -scale for magnons to heat up after photoexcitation, /em m emCg , decreases monotonically with increasing x, and does display structure near x ~ 0.25. Several theoretical models predict a strong correlation between τ D and αint. For example, Koopmans et al. predicts τ D will be inversely proportional to α by assuming that the dissipative processes responsible for damping also drive ultrafast demagnetization 27. Alternatively, Fähnle et al. predict s that τD should be proportional to αint 28. In our experiments on Co -Fe thin films, w e observe only a weak correlation between τD and αint. While α int varies with composition by a factor of three , τD for 8 of the 9 compositions we study fall within 20% of 75 fs. The τD value we obtained for Fe (= 76 fs) agrees well with experimental results reported in 9,12,29. 7 Conclusions We have measured the magnetization dynamics of Co xFe1-x thin-films , and we observe that both ultrafast and precessional dynamics of Co 0.25Fe0.75 differ significantly from Co and Fe . When the moment of Co0.25Fe0.75 is driven away from its equilibrium orientation , the time -scale for the moment to return to equilibrium is 3 -4x as long as for Fe or Co. Similarly, when spins of Co0.25Fe0.75 are driven into a nonequilibrium state by ultrafast laser heating, the time -scale for thermalization with the lattice is 2 -3x as long as for Fe or Co. Through 3TM analyses, we demonstrate that this occurs primarily due to the effect of the electronic band -structure on electron -phonon interactions , consistent with the “breathing Fermi surface” theory . Our findings are of fundamental importance to the field of ul trafast magnetism, which seeks to control magnetic order on femto - to picosecond time-scale s. Such control requires a thorough understanding of how and why energy is exchanged between electronic, spin, and vibrational degrees of freedom. Prior studies have shown that 𝑔𝑒𝑝 is correlated with a wide range of physical properties, e.g the superconducting transition temperature30, electrical resistivity 26, photoelectron emission31, and the laser fluence required for ablation32. To our knowledge, o ur study provides the first demonstration that 𝑔𝑒𝑝 in ferromagnetic metals is also correlated to the Gilbert damping parameter 𝛼. Our findings also have implications for the ongoing search for magnetic materials with ultrafast magnetic switching functionality. Atomistic spin dynamics simulations predict that the energy required for ultrafast electrical or optical switching of rare -earth ferromagnetic alloys, e.g. GdFeCo, is governed by the electron -phonon energy transfer coefficient33. To date, most studies aimed at exploring the materials science of ultrafast switching have used alloy composition as a way to control magnetic properties 34–37. Our work suggests an alternative strategy for reducing the energy requirements for ultrafast magnetic switching. The alloy composition should be chosen to minimize the electronic density of states at the Fermi -level. Such metals will have lower electron -phonon energy trans fer coefficients, and therefore more energy efficient ultrafast switching 33. Finally , our findings offer a new route for discovering ferro magnetic materials with ultra -low damping as a result of low 𝑔𝑒𝑝. Current methods for identifying low damping materials involve labor -intensive ferromagnetic resonance measurements of one alloy composition at a time. Alternatively, high-throughput localized measurements of ultrafast demagnetization dynamics of samples produced using combinatorial techniques38 would allow promising alloy compounds with weak electron -phonon interactions to be rapidly identified 39–41. 8 Materials and Methods Sample Preparation We sputter deposit the Co -Fe samples onto sapphire substrates with a direct current (DC) magnetron sputtering system (Orion, AJA International). The base pressure prior to deposition is less than 3.5 × 10-7 torr. We sputter with an Ar gon pressure of ~3.5 × 10-3 torr. The geometry of the samples is sapphire/Ta(2nm)/Cu(3nm)/Co xFe1-x(15nm)/Cu(3nm)/Ta(1nm). The Co xFe1-x layer is deposited by co -sputtering two 4N purity Co and Fe targets at different powers. We chose this film geometry to mimic the samples in Ref.6 which demonstrated low damping at x = 0.25. To ensure an accurate thickness of each layer in our samples, we calibrate the deposition rates of each metal by sputtering individual Co, Fe, Ta, and Cu films onto SiO 2/Si substrates and/or BK -7 glass substrates. We use picosecond acoustics42 and time-domain thermo -reflectance (TDTR) measurements43,44 to determine the thicknesses of these individual films. We validate the composition of the Co -Fe alloy layer by perf orming Energy Dispersive X -Ray Spectroscopy (EDS) analyses with a scanning electron microscope ( FEI Nova Nano SEM 450) at an operating voltage of 15 kV and working distance of 14 mm. We analyze the EDS data using Aztec Synergy software ( Oxford Instruments ). Time -Resolved MOKE Experimental Setup We use a pump/probe laser system to perform TR -MOKE measurements of the magne tization dynamics. The pulsed laser is a Ti:sapphire oscillator with an 80 MHz repetition rate. The laser beam is split into a pump and probe beam, that are modulated to frequencies of 10.7 MHz and 200 Hz , respectively. A time -delayed pump beam irradiates the sample surface and heats the metal film. The ultrafast heating causes a change in the magnetic moment. We measure the time - evolution of the magnetic moment by monitoring the polarization of the probe beam reflected of f the sample surface. The reflected probe beam’s polarization state is affected by the out -of-plane magnetic moment of the sample due to the polar Kerr effect. Additional details about the MOKE experiment set -up are in Ref.45. The t ime-resolution of our experiment is controlled by the convolution of the intensity vs. time of the pump and probe pulses. The wavelength of our pump and probe beams is tunable. Employing a red (900 nm ) pump and blue (450 nm ) probe yields higher time-resolution capabilities , allowing us to accurately measure the ultrafast magnetization at fe mtosecond time delays . We measure the full-width -at-half-maximum ( FWHM ) of the convolution of the pump and probe pulses by performing an inverse Faraday effect (IFE) measurement on Pt . We obtain a FWHM value of 390 fs for the convoluted pulses , and a pulse duration of 2 10 fs for the 900 nm pump/450 nm probe beam setup . For further details on our IFE measurements and pulse duration calculations, please refer to Supplementary Figure 8. 9 To investigate the precessional dynami cs on longer time -scales, we use a pump and probe wavelength of 783 nm. The pulse duration for this setup is 610 fs due to pulse broadening from a two-tint setup we use to prevent pump light from reaching the balanced detector45,46. Three Temperature Modeling To determine the electron, phonon, and magnon energy transfer coefficients, we use t he phenomenological three -temperature model (3TM), given by the following set of equations : 𝐶𝑒𝑑𝑇𝑒 𝑑𝑡=𝑔𝑒𝑝(𝑇𝑝−𝑇𝑒)+ 𝑔𝑒𝑚(𝑇𝑚−𝑇𝑒)+Ʌ𝑒𝑑2𝑇𝑒 𝑑𝑧2+𝑆(𝑧,𝑡) (1) 𝐶𝑝𝑑𝑇𝑝 𝑑𝑡=𝑔𝑒𝑝(𝑇𝑝−𝑇𝑒)+Ʌ𝑝𝑑2𝑇𝑝 𝑑𝑧2 (2) 𝐶𝑚𝑑𝑇𝑚 𝑑𝑡=𝑔𝑒𝑚(𝑇𝑚−𝑇𝑒)+Ʌ𝑚𝑑2𝑇𝑚 𝑑𝑧2 (3) 𝑆(𝑧,𝑡)= 𝑆0𝑃(𝑡)𝐴(𝑧) (4) Equations 1 – 3 describe the temperature evolution of electrons (e), phonons (p) and magnons (m), as a function of time delay (t). C, T, and Ʌ are the heat capacity per unit volume, temperature, and thermal conductivity, respectively. We use the density of states (DOS) at the Fermi level as a function of Co -concentration6 to calculate the electronic heat capacity (C e) using the Sommerfeld model . We assume that the phonon -magnon energy transfer is negligible compared to electron -magnon coupling, and thus, neglect 𝑔𝑝𝑚. We calculate the laser energy absorption by electrons (S), as a function of depth (z) and time delay (t), as described in Equation 4. The terms P(t) and A(z) denote the time -dependent laser pulse intensity and the optical absorption profile as a function of stack thickness. We calculat e A(z) us ing the refractive indices of each metal constituent of the stack47–49. The material parameters that are used to numerically solve equations 1 – 4 are listed in Supplementary Table 1. 10 Figures: Figure 1. Ultrafast magnetization dynamics of Co, Fe, and Co 0.25Fe0.75 thin films (a) Polar TR - MOKE data showing ultrafast demagnetization behavior at short delay times. (b) Schematic illustration of the three phases of an ultrafast magnetization dynamics experiment. Stage I: A large external magnetic field oriented normal to the plane of t he sample leads to an equilibrium moment , 𝑀⃗⃗ in the out -of-plane direction. Stage II: Upon heating with a pump beam, ultrafast demagnetization ( 𝑀′⃗⃗⃗⃗ ) occurs within ~100s of fs. Energy from hot electrons is transferred to the magnons, increasing the amplitude of precession. Stage III: Over the next few picoseconds, energy is transferred from magnons and electrons to the lattice. Additionally, spatial t emperature gradients relax. As a result, magnons cool, i.e. the average precessional amplitude of individual spins decreases. As a result, the magnetization partially recovers to 𝑀′′⃗⃗⃗⃗⃗⃗ . The time -scale for the partial recovery in stage III depends strongly o n the composition. 11 Figure 2 . Precessional dynamics in Co, Fe, and Co 0.25Fe0.75 thin films (a) Polar TR -MOKE data on sub -nanosecond time-scale s. (b) Illustration of the three stages for precessional dynamics after laser excitation . Stage I: Prior to laser excitation, the presence of a canted external magnetic field, 𝐻𝑒𝑓𝑓⃗⃗⃗⃗⃗⃗⃗⃗⃗ , oriented at an angle θ. This results in the orientation of the out -of-plane moments, 𝑀⃗⃗ 𝑧. Stage II: Laser -induced photoexcitation leads to the disorder of the magnetic moment, causing a decay in the net magnetization , denoted by 𝑀′⃗⃗⃗⃗ . The net torque imba lance causes macroscopic precessions of the magnons, towards equilibrium, 𝐻′𝑒𝑓𝑓⃗⃗⃗⃗⃗⃗⃗⃗⃗⃗ , over several ~100s of picoseconds . Stage III : Eventually, after ~1 ns, the magnetic moment re -equilibrates to 𝐻′𝑒𝑓𝑓⃗⃗⃗⃗⃗⃗⃗⃗⃗⃗ . The lifetime of the magnetic precessions depends o n the effective damping parameter, α eff. The time -scale for the precessional dynamics to cease ( in stage III) depends strongly on composition, and is a maximum for x = 0.25. 12 Figure 3. Compositional dependence of descriptors for the ultrafast dynamics data . (a) R describes the maximum change in the magnetic moment, i.e. how far from equilibrium spin - degrees of freedom are driven after ultrafast excitation. τD describes the lag between zero delay time and demagnetization, as a function of Co -concentration. (b) α denotes the Gilbert damping parameter, as a function of Co concentration. Data obtained from our TR -MOKE experiments described in this study (plotted in orange), agree reasonably with data from Ref. [6] (plotted in green). Co 0.25Fe0.75 features the largest deviation in R and α, when compared to its constituent elements Co and Fe. 13 Figure 4. Analyses of Ultrafast Demagnetization Results using the Three Temperature Model (3TM) in Co -Fe alloys . (a) Polar TR -MOKE dataset of the Co 0.25Fe0.75 composition (black circles) with best -fit results of the 3TM. The 3TM describes the temperature excursions of the electrons (blue curve), magnons (red curve) and phonons (green curve) after laser excitation. (b) We treat 𝑔𝑒𝑝 and 𝑔𝑒𝑚 as fit parameters when solving the 3TM. Using literature values of C p and C m (further details available in Supplementary Table 1), we calculate and plot the electron -phonon (τ ep) and electron -magnon (τ em) relaxation times, as a function of Co -concentration. The red -line is a best - fit value for the electron -phonon relaxation time as a function of composition, with the assumption of a composition -independent value for the electron -phonon coupling parameter λ . 14 References: 1. Kirilyuk, A., Kimel, A. V & Rasing, T. Ultrafast optical manipulation of magnetic order. Rev. Mod. Phys. 82, 2731 (2010). 2. Beaurepaire, E., Merle, J. C., Daunois, A. & Bigot, J. Y. Ultrafast spin dynamics in ferromagnetic nickel. Phys. Rev. Lett. 76, 4250 –4253 (1996). 3. Hellman, F. et al. Interface -Induced Phenomena in Magnetism. Rev. Mod. Phys. 89, 025006 (2017). 4. McMillan , W. L. Transition Temperature of Strong -Coupled Superconductors. Phys. Rev. 167, 331 –344 (1968). 5. Allen, P. B. Theory of thermal relaxation of electrons in metals. Phys. Rev. Lett. 59, 1460 – 1463 (1987). 6. Schoen, M. A. W. et al. Ultra -low magnetic damping of a metallic ferromagnet. Nat. Phys. 12, 839 –842 (2016). 7. Wei, Y. et al. Ultralow magnetic damping of a common metallic ferromagnetic film. Sci. Adv. 7, 1–7 (2021). 8. Lee, A. J. et al. Metallic ferromagnetic films with magne tic damping under 1.4 × 10 -3. Nat. Commun. 8, 1–6 (2017). 9. Koopmans, B. et al. Explaining the paradoxical diversity of ultrafast laser -induced demagnetization. Nat. Mater. 9, 259 –265 (2010). 10. Chen, Z. & Wang, L. W. Role of initial magnetic disorder: A time-dependent ab initio study of ultrafast demagnetization mechanisms. Sci. Adv. 5, eaau8000 (2019). 11. Carva, K., Battiato, M. & Oppeneer, P. M. Ab initio investigation of the Elliott -Yafet electron -phonon mechanism in laser -induced ultrafast demagneti zation. Phys. Rev. Lett. 107, 207201 (2011). 12. Carpene, E. et al. Dynamics of electron -magnon interaction and ultrafast demagnetization in thin iron films. Phys. Rev. B - Condens. Matter Mater. Phys. 78, 1–6 (2008). 13. Eich, S. et al. Band structure evo lution during the ultrafast ferromagnetic -paramagnetic phase transition in cobalt. Sci. Adv. 3, 1–9 (2017). 14. Carpene, E., Hedayat, H., Boschini, F. & Dallera, C. Ultrafast demagnetization of metals: Collapsed exchange versus collective excitations. Phys. Rev. B - Condens. Matter Mater. Phys. 91, 1–8 (2015). 15. Tengdin, P. et al. Critical behavior within 20 fs dr ives the out -of-equilibrium laser - induced magnetic phase transition in nickel. Sci. Adv. 4, 1–9 (2018). 16. Kimling, J. et al. Ultrafast demagnetization of FePt:Cu thin films and the role of magnetic heat capacity. Phys. Rev. B - Condens. Matter Mater. Phy s. 90, 1–9 (2014). 17. Wilson, R. B. & Coh, S. Parametric dependence of hot electron relaxation time-scale s on electron -electron and electron -phonon interaction strengths. Commun. Phys. 3, (2020). 15 18. Gilmore, K., Idzerda, Y. U. & Stiles, M. D. Identificat ion of the dominant precession - damping mechanism in Fe, Co, and Ni by first -principles calculations. Phys. Rev. Lett. 99, 1–4 (2007). 19. Kamberský, V. On the Landau –Lifshitz relaxation in ferromagnetic metals. Can. J. Phys. 48, 2906 –2911 (1970). 20. Haag, M., Illg, C. & Fähnle, M. Role of electron -magnon scatterings in ultrafast demagnetization. Phys. Rev. B - Condens. Matter Mater. Phys. 90, 1–6 (2014). 21. Tveten, E. G., Brataas, A. & Tserkovnyak, Y. Electron -magnon scattering in magnetic heterostructure s far out of equilibrium. Phys. Rev. B - Condens. Matter Mater. Phys. 92, 1–5 (2015). 22. Farle, M. Ferromagnetic resonance of ultrathin metallic layers. Reports Prog. Phys. 61, 755–826 (1998). 23. Schoen, M. A. W. et al. Magnetic properties in ultrathin 3d transition -metal binary alloys. II. Experimental verification of quantitative theories of damping and spin pumping. Phys. Rev. B 95, 1–9 (2017). 24. Kuneš, J. & Kamberský, V. First -principles investigation of the damping of fast magnetization precession in ferromagnetic (formula presented) metals. Phys. Rev. B - Condens. Matter Mater. Phys. 65, 1–3 (2002). 25. Fähnle, M. & Steiauf, D. Breathing Fermi s urface model for noncollinear magnetization: A generalization of the Gilbert equation. Phys. Rev. B - Condens. Matter Mater. Phys. 73, 1–5 (2006). 26. Allen, P. B. Empirical electron -phonon values from resistivity of cubic metallic elements. Phys. Rev. B 36, 2920 –2923 (1987). 27. Koopmans, B., Ruigrok, J. J. M., Dalla Longa, F. & De Jonge, W. J. M. Unifying ultrafast magnetization dynamics. Phys. Rev. Lett. 95, 1–4 (2005). 28. Zhang, W. et al. Unifying ultrafast demagnetization and intrinsic Gilbert damping in Co/Ni bilayers with electronic relaxation near the Fermi surface. Phys. Rev. B 96, 1–7 (2017). 29. Mathias, S. et al. Probing the time-scale of the exchange interaction in a ferromagnetic alloy. Proc. Natl. Acad. Sci. U. S. A. 109, 4792 –4797 (2012). 30. Brorson, S. D. et al. Femtosecond room -temperature measurement of the electron -phonon coupling constant in metallic superconductors. Phys. Rev. Lett. 64, 2172 –2175 (1990). 31. Gloskovskii, A. et al. Electron emission from films of Ag and Au nanoparticles excited by a femtosecond pump -probe laser. Phys. Rev. B - Condens. Matter Mater. Phys. 77, 1–11 (2008). 32. Chan, W. L., Averback, R. S., Cahill, D. G. & Lagoutchev, A. Dynamics of femtosecond laser -induced melting of silver. Phys. Rev. B - Condens. Matter Mater. Phys. 78, 1–8 (2008). 16 33. Atxitia, U., Ostler, T. A., Chantrell, R. W. & Chubykalo -Fesenko, O. Optim al electron, phonon, and magnetic characteristics for low energy thermally induced magnetization switching. Appl. Phys. Lett. 107, (2015). 34. Jakobs, F. et al. Unifying femtosecond and picosecond single -pulse magnetic switching in Gd-Fe-Co. Phys. Rev. B 103, 18–22 (2021). 35. Davies, C. S. et al. Pathways for Single -Shot All -Optical Switching of Magnetization in Ferrimagnets. Phys. Rev. Appl. 13, 1 (2020). 36. Ostler, T. A. et al. Ultrafast heating as a sufficient stimulus for magnetization reversal in a ferrimagnet. Nat. Commun. 3, (2012). 37. Ceballos, A. et al. Role of element -specific damping in ultrafast, helicity -independent, all - optical switching dynamics in amorphous (Gd,Tb)Co thin films. Phys. Rev. B 103, 24438 (2021). 38. Maier, W. F., Stowe, K. & Sieg, S. Combinatorial and high -throughput materials science. Angew. Chemie - Int. Ed. 46, 6016 –6067 (2007). 39. Geng, J., Nlebedim, I. C., Besser, M. F., Simsek, E. & Ott, R. T. Bulk Combinatorial Synthesis and High Throughput Characterization for Rapid Assessment of Magnetic Materials: Application of Laser Engineered Net Shaping (LENSTM). JOM 68, 1972 –1977 (2016). 40. Koinuma, H. & Takeuchi, I. Combinatorial solid -state chemistry of inorganic materials. Nature Materials 3, 429 –438 (2004). 41. Takeuchi, I., Lauterbach, J. & Fasolka, M. J. Combinatorial materials synthesis. Mater. Today 8, 18–26 (2005). 42. Hohensee, G. T ., Hsieh, W. P., Losego, M. D. & Cahill, D. G. Interpreting picosecond acoustics in the case of low interface stiffness. Rev. Sci. Instrum. 83, (2012). 43. Cahill, D. G. Analysis of heat flow in layered structures for time -domain thermoreflectance. Rev. Sc i. Instrum. 75, 5119 –5122 (2004). 44. Jiang, P., Qian, X. & Yang, R. Tutorial: Time -domain thermoreflectance (TDTR) for thermal property characterization of bulk and thin film materials. J. Appl. Phys. 124, (2018). 45. Gomez, M. J., Liu, K., Lee, J. G. & W ilson, R. B. High sensitivity pump -probe measurements of magnetic, thermal, and acoustic phenomena with a spectrally tunable oscillator. Rev. Sci. Instrum. 91, (2020). 46. Kang, K., Koh, Y. K., Chiritescu, C., Zheng, X. & Cahill, D. G. Two -tint pump -probe measurements using a femtosecond laser oscillator and sharp -edged optical filters. Rev. Sci. Instrum. 79, (2008). 47. Johnson, P. B. & Christy, R. W. Optical constants of transition metals. Phys. Rev. B 9, 5056 –5070 (1974). 48. P. B. Johnson and R. W. Chri sty. Optical Constant of the Nobel Metals. Phys. Rev. B 6, 17 4370 –4379 (1972). 49. Ordal, M. A., Bell, R. J., Alexander, R. W., Newquist, L. A. & Querry, M. R. Optical properties of Al, Fe, Ti, Ta, W, and Mo at submillimeter wavelengths. Appl. Opt. 27, 1203 (1988). Acknowledgements The work by R. M., V. H. O, and R. B. W. was primarily supported by the U.S. Army Research Laboratory and the U.S. Army Research Office under contract/grant number W911NF -18-1- 0364 and W911NF -20-1-0274. R. M. and R. B. W. also acknowledge support by NSF (C BET – 1847632). The work by L. V. and S. C. was supported by the U.S. Army Research Laboratory and U.S. Army Research Office under contract/grant number W911NF -20-1-0274. Energy Dispersive X -Ray Spectroscopy (EDS) analyses were performed at the Central Fac ility for Advanced Microscopy and Microanalysis (CFAMM) at UC Riverside. Author Contributions R. M. and R. B. W. designed the experiments. R. M. prepared all the samples and characterized them , and performed TR-MOKE experiments . V. H. O performed VSM measurements. L. V. performed hierarchical clustering analyses. S. C. performed DFT calculations. R. M. and R . B. W. analyzed the data and wrote the manuscript, with discussions and contributions from L. V. and S. C . Additional Information: Supplementary information is provided with this manuscript. Competing Interests: The authors declare no competing interest. Data Availability: The data that supports the findings of this paper are available from the corresponding author upon reasonable request. Correspondence: Correspondence and request for additional information must be addressed to rwilson@ucr.edu | 2021-07-24 | Recent investigations have advanced the understanding of how
structure-property relationships in ferromagnetic metal alloys affect the
magnetization dynamics on nanosecond time-scales. A similar understanding for
magnetization dynamics on femto- to pico-second time-scales does not yet exist.
To address this, we perform time-resolved magneto optic Kerr effect (TRMOKE)
measurements of magnetization dynamics in Co-Fe alloys on femto- to nano-second
regimes. We show that Co-Fe compositions that exhibit low Gilbert damping
parameters also feature prolonged ultrafast demagnetization upon
photoexcitation. We analyze our experimental TR-MOKE data with the
three-temperature-model (3TM) and the Landau-Lifshitz-Gilbert equation. These
analyses reveal a strong compositional dependence of the dynamics across all
time-scales on the strength of electron-phonon interactions. Our findings are
beneficial to the spintronics and magnonics community, and will aid in the
quest for energy-efficient magnetic storage applications. | Electron-Phonon Scattering governs both Ultrafast and Precessional Magnetization Dynamics in Co-Fe Alloys | 2107.11699v1 |
arXiv:2206.10948v1 [math.AP] 22 Jun 2022HOMOGENIZATION OF THE LANDAU-LIFSHITZ-GILBERT EQUATION WITH NATURAL BOUNDARY CONDITION JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN Abstract The full Landau-Lifshitz-Gilbert equation with periodic material coe fficients and natural boundary condition is employed to model the magnetization dynamics in composite ferromagnets. In this work, we establish the converge nce between the homogenized solution and the original solution via a Lax equivalence th eorem kind of argument. There are a few technical difficulties, including: 1) it is p roven the classic choice of corrector to homogenization cannot provide the c onvergence re- sult in the H1norm; 2) a boundary layer is induced due to the natural boundary condition; 3) the presence of stray field give rise to a multiscale pote ntial problem. To keep the convergence rates near the boundary, we introduce the Neumann cor- rector with a high-order modification. Estimates on singular integra l for disturbed functions and boundary layer are deduced, to conduct consisten cy analysis of stray field. Furthermore, inspired by length conservation of magnetizat ion, we choose proper correctors in specific geometric space. These, together with a uniform W1,6 estimate on original solution, provide the convergence rates in the H1sense. 1.Introduction The intrinsic magnetic order of a rigid single-crystal ferr omagnet over a region Ω ⊂Rn,n= 1,2,3 is described by the magnetization Msatisfying M=Ms(T)m,a.e. in Ω , where the saturation magnetization Msdepends on the material and the temperature T. Below Curie temperature, Msis modeled as a constant. A stable structure of a ferromagnet is mathematically chara cterized as the local minimizers of the Landau-Lifshitz energy functio nal [7] GL[m] :=/integraldisplay Ωa(x)|∇m|2dx+/integraldisplay ΩK(x)(m·u)2(m)dx −µ0/integraldisplay Ωhd[Msm]·Msmdx−/integraldisplay Ωha·Msmdx =:E(m)+A(m)+W(m)+Z(m). Date: June 23, 2022. 2010Mathematics Subject Classification. 35B27; 65M15; 82D40. Key words and phrases. Homogenization; Landau-Lifshitz-Gilbert equation; Boun dary layer; Magnetization dynamics; Micromagnetics. 12 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN E(m) is the exchange energy, which penalizes the spatial variat ion ofm. The matrix a= (aij)1≤i,j≤3is symmetric, uniformly coercive and bounded, i.e., (1) n/summationdisplay i,j=1aij(x)ηiηj≥amin|η|2for anyx∈Rn,η∈Rn, n/summationdisplay i,j=1aij(x)ηiξj≤amax|η||ξ|for anyx∈Rn,η, ξ∈Rn. Intheanisotropyenergy A(m),uistheeasy-axisdirectionwhichdependson thecrystallographicstructureofthematerial. Theanisot ropyenergydensity is assumed to be a non-negatively even and globally Lipschit z continuous function that vanishes only on a finite set of unit vectors (th e easy axis). The third term W(m) is the magnetostatic self-energy due to the dipolar magnetic field, also known as the stray field hd[m]. For an open bounded domain Ω with a Lipschitz boundary, the magnetization m∈Lp(Ω,R3) generates a stray field satisfying (2) hd[m] =∇Um, where the potential Umsolves (3) ∆ Um=−div(mXΩ),inD′(R3) withmXΩthe extension of mtoR3that vanishes outside Ω. The exis- tence and uniqueness of Umfollows from the Lax-Milgram Theorem and Um satisfies the estimate [ 10] (4) /ba∇dblhd[m]/ba∇dblLp(Ω)≤ /ba∇dblm/ba∇dblLp(Ω)1< p <∞. The last term Z(m) is the Zeeman energy that models the interaction be- tweenmand the externally applied magnetic field ha. For a composite ferromagnet with periodic micorstructures , the material constants are modeled with periodic material coefficients wi th period ε, i.e., aε=a(x/ε),Kε=K(x/ε),Mε=Ms(x/ε), with functions a,K,Ms periodic over Y= [0,1]n. The associated energy reads as (5)Gε L[m] :=/integraldisplay Ωaε(x)|∇m|2dx+/integraldisplay ΩKε(x)(m·u)2dx −µ0/integraldisplay Ωhd[Mεm]·Mεmdx−/integraldisplay Ωha·Mεmdx. It is proved in [ 2] thatGε L[m] is equi-mild coercive in the metric space (H1(Ω,S2),dL2(Ω,S2)) and Γ-converges to the functional Ghomdefined as Ghom[m] =/integraldisplay Ωa0|∇m|2dx+/integraldisplay ΩK0(m·u)2dx−µ0(M0)2/integraldisplay Ωhd[m]·mdx −µ0/integraldisplay Ω×Y/vextendsingle/vextendsinglem·Hd[Ms(y)](y)/vextendsingle/vextendsingle2dxdy−M0/integraldisplay Ωha·mdx, (6)HOMOGENIZATION OF THE LLG EQUATION 3 wherea0is the homogenized tensor a0 ij=/integraldisplay Y/parenleftBigg aij+n/summationdisplay k=1aik∂χj ∂yk/parenrightBigg dy, the constants M0andK0are calculated by M0=/integraldisplay YMs(y)dy, K0=/integraldisplay YK(y)dy, and the symmetric matrix-valued function Hd[Ms(y)](y) =∇yU(y) with potential function given by (7)/integraldisplay YMs(y)∇yϕ(y)dy=−/integraldisplay Y∇yU(y)·∇yϕ(y)dy, U(y) isY-periodic ,/integraldisplay YU(y)dy= 0, for any periodic function ϕ∈H1 per(Y). In the current work, we are interested in the convergence of t he dynamic problem driven by the Landau-Lifshitz energy ( 5) to the dynamics problem driven by the homogenized energy ( 6) asεgoes to 0. It is well known that the time evolution of the magnetization over Ω T= Ω×[0,T] follows the Landau-Lifshitz-Gilbert (LLG) equation [ 7,6] (8) ∂tmε−αmε×∂tmε=−(1+α2)mε×Hε e(mε) a.e. in Ω T, ν·aε∇mε= 0,a.e. on∂Ω×[0,T], mε(0,x) =mε init(x),|mε init(x)|= 1 a.e. in Ω , whereα >0isthedampingconstant, andtheeffective field Hε e(mε) =−δGε L δmε associated to the Landau-Lifshitz energy ( 5) is given by (9)Hε e(mε) = div(aε∇mε)−Kε(mε·u)u+µ0Mεhd[Mεmε]+Mεha. Meanwhile, the LLG equation associated to the homogenized e nergy (6) reads as (10) ∂tm0−αm0×∂tm0=−(1+α2)m0×H0 e(m0) ν·a0∇m0= 0,a.e. on∂Ω×[0,T] m0(0,x) =m0 init(x),|minit(x)|= 1 a.e. in Ω with homogenized effective field H0 e(m0) =−δGhom δm0calculated by (11)H0 e(m0) =div/parenleftbig a0∇m0/parenrightbig −K0(m0·u)u +µ0(M0)2hd[m0]+µ0H0 d·m0+M0ha, where the matrix H0 d=/integraltext YMs(y)Hd[Ms(y)](y)dy. Works related tothehomogenization oftheLLGequationinth eliterature include [ 11,5,1,8,9,4]. As for the convergence rate, most relevantly, the LLG equation ( 8) with only the exchange term and with the periodic4 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN boundarycondition is studied in [ 8]. Convergence rates between mεandm0 in time interval [0 ,εσT] are obtained under the assumption (12) /ba∇dbl∇mε/ba∇dblL∞(Ω)≤C,for anyt∈[0,εσT], whereCis a constant independent of εandσ∈[0,2). As a special case, whenσ= 0 in assumption ( 12), i.e.,/ba∇dbl∇mε/ba∇dblL∞(Ω)is uniformly boundedover a time interval independent of ε, it is proven that /ba∇dblmε−m0/ba∇dblL∞(0,T;L2(Ω))= O(ε)while/ba∇dblmε−m0/ba∇dblL∞(0,T;H1(Ω))isonlyuniformlyboundedwithoutstrong convergence rate. In this work, we consider the full LLG model ( 8) equipped with the Neu- mann boundary condition, which is the original model derive d by Landau and Lifshitz [ 7]. We prove the convergence rates between mεandm0in the L∞(0,T;H1(Ω)) sense without the strong assumption ( 12). It is worth men- tioning that, the trick to improve the convergence result in toH1sense is to find proper correctors m1,m2, such that they satisfy geometric properties (13) m0·m1= 0,andm0·m2=−|m1|2, which are motivated by the length-preserving property of ma gnetization and asymptotic expansion. A familiar definition of classic fi rst-order ho- mogenization corrector m1in (32) would naturally satisfies first property in (13); see [8]. In this article, the suitable corrector m2in (13) is obtained. By the usage of these properties, we are able to derive the estim ate of consis- tency error, which is induced by an equivalent form of LLG equ ation, given in (22), and a sharper estimate than [ 8] inL∞(0,T;H1(Ω)) sense is finally obtained. Instead of the assumption ( 12), we prove a weak result that /ba∇dbl∇mε/ba∇dblL6(Ω) is uniformly bounded over a time interval independent of ε. Such a uniform estimate is nontrivial for the LLG equation, since the stand ard energy es- timate usually transforms the degenerate (damping) term in to the diffusion term and thus the upper bound becomes ε-dependent. To overcome this difficulty, we introduce the interpolation inequality when n≤3 (14)/ba∇dbldiv(aε∇m)/ba∇dbl3 L3(Ω)≤C+C/ba∇dbldiv(aε∇m)/ba∇dbl6 L2(Ω) +C/ba∇dblm×∇{div(aε∇m)}/ba∇dbl2 L2(Ω), for theS2-value function msatisfying homogeneous Neumann boundary condition. This inequality can help us derive a structure-p reserving energy estimate, in which the degenerate term is kept in the energy. The full LLG model ( 8) we considered contains the stray field, where an independent homogenization problem of potential function in the distribu- tion sense arises, and this complicated the problem when we a rrive at the consistency analysis. By using results in [ 10] and Green’s representation formula, the stray field is rewritten as the derivatives of Ne wtonian poten- tial. Then we are able to obtain the consistency error by deri ving detailed estimate of singular integral for disturbed function and bo undary layer.HOMOGENIZATION OF THE LLG EQUATION 5 The effect of boundary layer exists when we apply classic homog eniza- tion corrector to the Neumann boundary problem, which would cause the approximation deterioration on the boundary. To avoid this , a Neumann corrector is introduced, which is usually used in elliptic h omogenization problems (see [ 12] for example). In this article, we provide a strategy to apply the Neumann corrector to evolutionary LLG equations, by finding a proper higher-order modification. For a big picture, let us w rite ahead the linear parabolic equation of error ∂teε−Lεeε+fε=0, whose detailed derivation can be found in ( 26). Following the notation of eε b=eε−ωbwith boundary corrector ωb, one can find by above equa- tion that an L∞(0,T;H1(Ω)) norm of eε brelies on the boundary data and inhomogeneous term induced by ωb, which read as (15) /ba∇dblν·aε∇{eε+ωb}/ba∇dblB−1/2,2(∂Ω)and/ba∇dblLεωb/ba∇dblL2(Ω). In this end, we divide the corrector ωbinto two parts as ωb=ωN−ωM, such that they can control two terms in ( 15) respectively. Here ωNis the Neumann corrector used in elliptic problems (see [ 12]), andωMis a modi- fication to be determined. We point out the modification ωMis necessary since calculation implies some bad terms in LεωNdo not converge in L2 sense. Therefore we construct following elliptic problem t o determine ωM: div(aε∇ωM) =/parenleftBig Bad Terms in LεωN/parenrightBig with proper Neumann boundary condition. Such a solution ωMcan be proved to have better estimates than ωN, by the observation that all “Bad Terms in LεωN” can be written in the divergence form. At this point, ωM can be viewed as a high-order modification. This paper is organized as follows. In the next Section, we in troduce the main result of our article and outline the main steps of the pr oof. In Sec- tion3, multiscale expansions are used to derive the second-order corrector m2. In Section 4, we deduce that the consistency error fεonly relies on the consistency error of the stray field, which can be estimated b y calculation of singular integral for disturbed function and boundary la yer. In Section 5, we introduce the boundary corrector ωb, and derive several relevant es- timates of it. Section 6contains the stability analysis in L2andH1sense respectively. And we finally give a uniform regularity analy sis ofmε, by deriving a structure-preserving energy estimate in Sectio n7. 2.Main result To proceed, we make the following assumption Assumption 1. 1.Smoothness We assume Y-periodic functions a(y) = (aij(y))1≤i,j≤3, K(y),Ms(y), and the time-independent external field ha(x), alone with6 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN boundary ∂Ω, are sufficiently smooth. These together with the definition in (32),(46),(66),(76)leads to the smoothness of m1,m2andωb. 2.Initial data Assumem0 init(x)andmε init(x)are smooth enough and satisfy the Neumann compatibility condition: (16) ν·aε∇mε init(x) =ν·a0∇m0 init(x) = 0,x∈∂Ω. Furthermore, we might as well set them satisfying periodica lly disturbed el- liptic problem: (17) div(aε∇mε init(x)) = div( a0∇m0 init(x)),x∈Ω. (16)-(17)implym0 init(x)is the homogenization of mε init(x). Note that the as- sumption (17)is necessary not only for the convergence analysis in Theorem 4, but also for the uniform estimate of mεin Theorem 8. Now let us state our main result: Theorem 1. Letmε∈L∞(0,T;H2(Ω)),m0∈L∞(0,T;H6(Ω))be the unique solutions of (8)and(10), respectively. Under Assumption 1, there exists some T∗∈(0,T]independent of ε, such that for any t∈(0,T∗)and forn= 2,3, it holds (18)/ba∇dblmε(t)−m0(t)/ba∇dblL2(Ω)≤β(ε),/ba∇dblmε(t)−m0(t)/ba∇dblH1(Ω)≤Cε1/2, where (19) β(ε) =/braceleftBigg Cε[ln(ε−1+1)]2,whenn= 2, Cε5/6,whenn= 3. In the absence of the stray field, i.e., µ0= 0, then it holds for any t∈(0,T∗) and forn= 1,2,3 (20)/ba∇dblmε(t)−m0(t)/ba∇dblL2(Ω)≤Cε[ln(ε−1+1)]2, /ba∇dblmε(t)−m0(t)−(Φ−x)∇m0(t)/ba∇dblH1(Ω)≤Cε[ln(ε−1+1)]2, wherexis spatial variable, Φ= (Φi)1≤i≤nis the corrector defined in (67). Constant Cdepends on the initial data mε initandm0 init, but is independent ofε. Remark 2.1. Comparing (18)and(20), one can see that in the L2norm, the stray field makes little influence when n= 2, but causes 1/6-order loss of rate when n= 3. In the H1norm, however, the stray field leads to 1/2- order loss of rate in both cases. Such a deterioration of conv ergence rate is induced since the zero-extension has been applied for stray field(3), which introduces a boundary layer. Remark 2.2. The logarithmic growth [ln(ε−1+1)]2in(20)is caused by the Neumann corrector (Φ−x)∇m0. For problems (8)and(10)with periodic boundary condition over a cube, by replacing the Neumann cor rector in (20)HOMOGENIZATION OF THE LLG EQUATION 7 with the classical two-scale corrector, a similar argument in the current work leads to (21) /ba∇dblmε−m0−χ∇m0/ba∇dblH1(Ω)≤Cε, whereχ= (χi)1≤i≤nis defined in (33). Note that (21)is consistent with the L2result in [8]. However, only the uniform boundedness in H1has been shown in [8], while our results (20)and (21)imply that it maintains the same convergence rate in L2andH1norm, by choosing the correctors satisfying specific geometric pr operty(13). 2.1.Some notations and Lax equivalence type theorem. Recall that a classical solution to ( 8) also satisfies an equivalent form of equation, reads (22)LLLG(mε) :=∂tmε−αHε e(mε)+mε×Hε e(mε)−αgε l[mε]mε= 0, where the gε l[·] is the energy density calculated by (23)gε l[mε] =aε|∇mε|2+Kε(mε·u)u−hd[Mεmε]·Mεmε−ha·Mεmε. For convenience, we also define a bilinear operator deduced f rom (23), which reads Bε[m,n] =aε∇m·∇n+Kε(m·u)(n·u)−µ0hd[Mεm]·Mεn. Now let us set up the equation of error, in terms of Lax equival ence theorem kind of argument. Define the approximate solution (24) /tildewidemε(x) =m0(x)+εm1(x,x ε)+ε2m2(x,x ε), wherem0is the homogenized solution to ( 10),m1is the first-order corrector definedin( 32), andm2isthesecond-ordercorrectordeterminedbyTheorem 2. Then replacing mεby/tildewidemεin (22) provides the notation of consistence errorfε: (25) LLLG(/tildewidemε) =fε. Together ( 22) and (25), we can obtain the equation of error eε=mε−/tildewidemε, denoted by (26) ∂teε−Lεeε+fε=0, whereLεis second-order linear elliptic operator depending on mεand/tildewidemε, that can be characterized as (27) Lε(eε) =α/tildewideHε e(eε)−D1(eε)−D2(eε). Here/tildewideHε e(mε) is the linear part of Hε e(mε), i.e., /tildewideHε e(mε) :=Hε e(mε)−Mεha, procession term D1is calculated by (28)D1(eε) =mε×Hε e(mε)−/tildewidemε×Hε e(/tildewidemε) =mε×/tildewideHε e(eε)+eε×Hε e(/tildewidemε),8 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN and the degeneracy term D2reads as D2(eε) =−αgε l[mε]mε+αgε l[/tildewidemε]/tildewidemε =−α/parenleftbig Bε[eε,mε]+Bε[/tildewidemε,eε]+Mε(ha·eε)/parenrightbig mε−αgε l[/tildewidemε]eε. Moreover, we define a correctional error eε bas (29) eε b=eε−ωb, whereωbis the boundary corrector satisfying ωb=ωN−ωM, forωNthe Neumann corrector given in ( 66), andωMthe modification determined in (76). Then equation ( 26) leads to (30) ∂teε b−Lεeε b+/parenleftbig ∂tωb−Lεωb+fε/parenrightbig = 0. By the Lax equivalence theorem kind of argument, the estimat e of error eε bfollows from consistency analysis of ( 25), energy estimate of boundary corrector, and stability analysis of ( 30). 2.2.Proof of Theorem 1. Proof.Following the above notations, for the consistency error fε, Theorem 3says that it can be divided as fε=f0+/tildewidef, satisfying /ba∇dbl/tildewidef(t)/ba∇dblL2(Ω)≤Cε, and /ba∇dblf0(t)/ba∇dblL2(Ω)= 0,whenµ0= 0, /ba∇dblf0(t)/ba∇dblLr(Ω)≤Crµ0/parenleftbig ε1/r+εln(ε−1+1)/parenrightbig ,whenµ0>0,n/ne}ationslash= 1, where constants CrandCare independent of ε, for any t∈(0,T), and 1≤r <+∞. Consideringtheboundarycorrectortermsin( 30), byTheorem 5there exists C=C(/ba∇dbl∇mε/ba∇dblL2(Ω)) such that /ba∇dbl∂tωb(t)/ba∇dblL2(Ω)≤Cεln(ε−1+1), /ba∇dblLεωb(t)/ba∇dblL2(Ω)≤Cε[ln(ε−1+1)]2+C/ba∇dbleε b(t)/ba∇dblH1(Ω), for anyt∈(0,T). As for initial-boundary data of eε b, using Theorem 4we write with C=C(/ba∇dbl∇mε/ba∇dblL2(Ω)), /ba∇dbleε b(x,0)/ba∇dblH1(Ω)+/ba∇dbl∂ ∂νεeε b/ba∇dblW1,∞(0,T;B−1/2,2(∂Ω))≤Cεln(ε−1+1). Now let us turn to stability analysis of ( 30). For the L∞(0,T;L2(Ω)) norm, let σ= 1 when n= 1,2, andσ= 6/5 whenn= 3, we can apply Theorem 6to derive for n= 1,2,3 (31)/ba∇dbleε b/ba∇dbl2 L∞(0,T;L2(Ω))+/ba∇dbl∇eε b/ba∇dbl2 L2(0,T;L2(Ω)) ≤Cδ/parenleftBig /ba∇dbleε b(x,0)/ba∇dbl2 L2(Ω)+/ba∇dbl∂ ∂νεeε b/ba∇dbl2 L2(0,T;B−1/2,2(∂Ω))+/ba∇dbl/tildewidef/ba∇dbl2 L2(0,T;L2(Ω)) +/ba∇dbl∂tωb/ba∇dbl2 L2(0,T;L2(Ω))+γ(ε)/ba∇dblf0/ba∇dbl2 L2(0,T;Lσ(Ω))/parenrightBig +δ/ba∇dblLεωb/ba∇dbl2 L2(0,T;L2(Ω))+ε2/ba∇dblAεeε b/ba∇dbl2 L2(0,T;L2(Ω)).HOMOGENIZATION OF THE LLG EQUATION 9 with /braceleftBigg γ(ε) = 1, whenn= 1,3, γ(ε) = [ln(ε−1+1)]2,whenn= 2. Constant Cδ=Cδ(/ba∇dbl∇mε/ba∇dblL4(Ω)). Now taking δsmall enough in ( 31), and using the fact /ba∇dblAεeε b/ba∇dblL2(0,T;L2(Ω))≤Cln(ε−1+1) withC=C/parenleftbig /ba∇dblAεmε/ba∇dblL2(Ω)/parenrightbig from Theorem 5, we finally obtain /ba∇dbleε b/ba∇dblL∞(0,T;L2(Ω))≤/braceleftBiggβ(ε), whenµ0>0,n= 2,3, Cε[ln(ε−1+1)]2,whenµ0= 0,n= 1,2,3, whereβ(ε) satisfies ( 19). Using the fact mε−m0=eε b+εm1+ε2m2+ωb, along with the estimates of εm1,ε2m2,ωbin Lemma 4-5, we obtain the L2 estimates in Theorem 1. As for the stability of ( 30) inL∞(0,T;H1(Ω)) norm, we can apply The- orem7to obtain for n= 1,2,3 /ba∇dbl∇eε b/ba∇dbl2 L∞(0,T;L2(Ω))≤C/parenleftBig /ba∇dbleε b(x,0)/ba∇dbl2 H1(Ω)+/ba∇dbl∂ ∂νεeε b/ba∇dbl2 H1(0,T;B−1/2,2(∂Ω)) +/ba∇dblLεωε b/ba∇dbl2 L2(0,T;L2(Ω))+/ba∇dblfε/ba∇dbl2 L2(0,T;L2(Ω))+/ba∇dbl∂tωb/ba∇dbl2 L2(0,T;L2(Ω))/parenrightBig , where constant C=C(/ba∇dblAεmε/ba∇dblL2(Ω),/ba∇dbl∇mε/ba∇dblL4(Ω)). Together with above results, and estimate for /ba∇dbl∇eε b/ba∇dbl2 L2(0,T;L2(Ω))in (31), we arrive at /ba∇dbl∇eε b(t)/ba∇dbl2 L∞(0,T;L2(Ω))≤/braceleftBigg Cε1/2, whenµ0>0,n= 2,3, Cε[ln(ε−1+1)]2,whenµ0= 0,n= 1,2,3, by the representation of eε bin (80), together with estimate of m2andωM in Lemma 5, it leads to the H1estimates in Theorem 1. Notice that all the constants in our estimate depend on the va lue of /ba∇dblAεmε(t)/ba∇dblL2(Ω)and/ba∇dbl∇mε(t)/ba∇dblL4(Ω), which from Theorem 8are uniformly bounded with respect to εandtfor anyt∈(0,T∗), with some T∗∈(0,T]. This completes the proof. /square 3.The Asymptotic Expansion In this section, we derive the second-order corrector using the formal asymptotic expansion. First, let us define the first-order co rrectorm1by (32) m1(x,y) =n/summationdisplay j=1χj(y)∂ ∂xjm0(x), whereχj,j= 1,...,nare auxiliary functions satisfying cell problem (33) div/parenleftbig a(y)∇χj(y)/parenrightbig =−n/summationdisplay i=1∂ ∂yiaij(y), χjY-periodic ,10 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN such that the first geometric property in ( 13) holds. As for the second-order corrector m2, we assume it as a two-scale function satisfying /braceleftBigg m2(x,y) is defined for x∈Ω andy∈Y, m2(·,y) isY-periodic . For notational convenience, given a two-scale function in t he form of m(x,x ε), we denote the fast variable y=x εand have the following chain rule (34) ∇m(x,x ε) = [(∇x+ε−1∇y)m](x,y). Moreover, denoting Aε= div(aε∇), one can rewrite Aεm(x,x ε) = [(ε−2A0+ε−1A1+A2)m](x,y), where (35) A0= divy/parenleftbig a(y)∇y/parenrightbig , A1= divx/parenleftbig a(y)∇y/parenrightbig +divy/parenleftbig A(y)∇x/parenrightbig , A2= divx/parenleftbig a(y)∇x/parenrightbig . The procedure to determine m2is standard. With the notation in ( 24), assumemεcan be written in form of (36) mε(x) =/tildewidemε(x)+o(ε2). One can derive m2by substituting ( 36) into (8) and comparing like terms of ε. However, it is a bit fussy in the presence of stray field. Let u s outline the main steps here. Revisiting the stray field hd[Mεmε(x)] =∇Uεin (2)-(3), one finds that the potential function Uε=Uε[Mεmε(x)] satisfies (37) ∆Uε=−div(Ms(x ε)mεXΩ). Substituting Uε= Σ2 j=0εjUj(x,x ε)+o(ε2) and (36) into (37) and combining like terms of εleads to (38) divy(∇yU0(x,y)) = 0, divy(∇yU1(x,y)) =−divy(Ms(y)m0(x)XΩ(x)), divx(∇xU0(x,y))+2div y(∇xU1(x,y))+div y(∇yU2(x,y)) =−Ms(y)divxm0(x)XΩ(x)−divy(Ms(y)m1(x,y)XΩ(x)). The first equation in ( 38) implies that U0(x,y) =U0(x) since the Lax- Milgram Theorem ensures the uniqueness and existence of sol ution (up to a constant). Integrating the third equation in ( 38) with respect to yyields ∆U0(x) =−div(M0m0XΩ). An application of ( 2)-(3) implies that U0is actually the potential function ofhd[Mhm0], i.e., (39) ∇U0(x) =hd[Mhm0] =Mhhd[m0].HOMOGENIZATION OF THE LLG EQUATION 11 With notation given in ( 7), one can deduce from the second equation in ( 38) thatm0(x)XΩ(x)U(y) =U1(x,y) up to a constant in the H1(Y) space. Hence it follows that by ( 7) (40) ∇yU1(x,y) =XΩ(x)m0(x)·Hd[Ms(y)](y). Substituting ( 39) and (40) into the expansion of Uε, one can deduce that, forx∈Ω, (41)hd[Mεmε] =∇Uε=hd[Mhm0]+m0(x)·Hd[Ms(y)](x ε)+O(ε). Substituting ( 36), (32), (35), (41) into (8) and collecting terms of O(ε0), we obtain the following equations (42)/braceleftBigg ∂tm0−αm0×∂tm0=−(1+α2)m0×{A0m2+Ha e}, m2Y-periodic in y, where (43)Ha e=A1m1+A2m0−Kε(m0·u)u +µ0Mshd[Mhm0]+µ0Msm0·Hd[Ms(y)]+Msha. Substituting ( 10) into (42) leads to (44)/braceleftBigg m0×A0m2=m0×/braceleftbig H0 e(m0)−Ha e/bracerightbig , m2Y-periodic in y. (44) is the degenerate system that determines m2in terms of m0. 3.1.Second-order corrector. Thewell-posednessof ( 44)isnontrivial due tothedegeneracy. InthefollowingTheorem, bysearchingas uitablesolution satisfying ( 13), we give the existence result, andderive an explicit expre ssion ofm2in terms of m0and some auxiliary functions. Theorem 2. Givenm0∈L∞/parenleftbig [0,T];H2(Ω)/parenrightbig the homogenization solution andm1calculated in (32), define V=/braceleftBig m∈H2(Y)∩H1 per(Y) :m·m0=−1 2|m1|2a.e. inΩ×Y/bracerightBig , then(44)admits a unique solution m2(x,y)∈ V/Tm0(S2), with notation Tm0(S2)denoting the tangent space of m0. Proof.Assume m2(x,y)∈ V, i.e.,m2·m0=−1 2|m1|2. Applying A0to both sides of it yields (45) m0·A0m2=−a(y)∇ym1·∇ym1−m1·A0m1. Taking the cross-product with m0to (44) and substituting ( 45) lead to (46)A0m2=−{Ha e−H0 e(m0)}+/braceleftbig m0·/parenleftbig Ha e−H0 e(m0)/parenrightbig/bracerightbig m0 −(m1·A0m1+a(y)∇ym1·∇ym1)m0.12 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN Now using the fact (47)/integraldisplay YH0 e(m0)−Ha edy= 0, together with ( 45), one can check equation ( 46) satisfies the compatibility condition for Y-periodic function m2iny. Thus by the application of Lax- Milgram Theorem andsmoothnessassumption, ( 46) admitsauniqueregular solution up to a function independent of y, denoted by m2(x,y)+/tildewidem2(x). Moreover, one can determine /tildewidem2(x) by m0·(m2+/tildewidem2) =−1 2|m1|2, such that m2+/tildewidem2∈ V, and therefore is also a solution to ( 44) by taking the above transformation inversely. /square Remark 3.1. One can check that equation (46)has a solution (48) m2=n/summationdisplay i,j=1θij∂2m0 ∂xi∂xj+n/summationdisplay i,j=1(θij+1 2χiχj)/parenleftbigg∂m0 ∂xi·∂m0 ∂xj/parenrightbigg m0+Tlow−(m0·Tlow)m0 with low-order terms Tlowcalculated by Tlow=−κ(m0·u)u+µ0ρhd[Mhm0]+µ0m0·Λ+Msha, whereθijandκ,ρ,Λare given by (49) A0θij=a0 ij−/parenleftbig aij+n/summationdisplay k=1aik∂χj ∂yk/parenrightbig −n/summationdisplay k=1∂(aikχj) ∂yk, A0ρ=Ms(y)−M0,A0κ=K(y)−K0, A0Λ=Ms(y)Hd[Ms(y)](y)−H0 d, θij, κ, ρ,Λ,areY-periodic . Moreover, one can find m2defined above satisfies geometric property (13), therefore is also the solution to equation (44). In the following, we may assume second-order correct m2takes the form in (48). 4.Consistency Estimate In this section, we aim to estimate the consistence error fεdefined in (25). Following the notation in ( 34)-(35), by the definition of /tildewidemε, (25) can be written in terms of fε=ε−2f−2+ε−1f−1+f0+εf1+ε2f2. It is easy to check that f−2=f−1=0by the definition of m0,m1in Section3. Along the same line, by the H¨ older’s inequality, one has /ba∇dblf1/ba∇dblL2(Ω)+/ba∇dblf2/ba∇dblL2(Ω)≤C,HOMOGENIZATION OF THE LLG EQUATION 13 whereCdepends on the L2(Ω) andL∞(Ω) norms of mi(x,x ε),∇xmi(x,x ε), ∇ymi(x,x ε),i= 0,1,2, and thus is bounded from above by /ba∇dbl∇m0/ba∇dblH4(Ω) with the help of smoothness assumption and Sobolev inequali ty. It remains to estimate f0, let us prove that f0only depends on the con- sistence error of stray field, by the help of geometric proper ty (13). Denote the consistence error of stray field by (50)/tildewideh=µMεhd[(Mε−Mh)m0]−µMεHd[Ms(y)](x ε)·m0 withHdgiven in ( 7). After some algebraic calculations and the usage of (42) and (43), one has (51)f0=∂tm0−α/braceleftBig A0m2+Ha e+/tildewideh/bracerightBig +m0×/braceleftBig A0m2+Ha e+/tildewideh/bracerightBig −αgε l[m0]m0−(aε∇ym1·∇ym1)m0−2(aε∇ym1·∇m0)m0. Notice that the classical solution m0to (10) also satisfies the equivalent form ∂tm0−αH0 e(m0)+m0×H0 e(m0)−αg0 l[m0]m0= 0, (52) where g0 l[m] :=a0|∇m|2+K0(m·u)u−µ0(M0)2hd[m]·m −µ0m·H0 d·m−ha·M0m. Substituting ( 52) into (51) and using ( 46) lead to f0=−α/tildewideh+m0×/tildewideh−α/braceleftbig m0·/parenleftbig Ha e−H0 e(m0)/parenrightbig/bracerightbig m0 +α(m1·A0m1−2aε∇ym1·∇m0)m0+αg0 l[m0]m0−αgε l[m0]m0. (53) Note that A2m0=Aεm0−ε−1A1m0, one can deduce Ha e=Hε e(m0)+A1m1+ε−1A0m1−/tildewideh. Substituting it into ( 53), and using the fact m0·Hε e(m0) =−gε l[m0],m0·H0 e(m0) =−g0 l[m0], one has (54)f0=−α/tildewideh+m0×/tildewideh−α/braceleftbig m0·/parenleftbig A1m1+ε−1A0m1−/tildewideh/parenrightbig/bracerightbig m0 +α(m1·A0m1−2aε∇ym1·∇m0)m0. ApplyA0andA1to both sides of m0·m1= 0 respectively, and substitute resulting equations into ( 54). After simplification, we finally obtain (55) f0=−α/tildewideh+m0×/tildewideh+α/parenleftBig m0·/tildewideh/parenrightBig m0. (55) implies that the convergence of fεdepends on the convergence of stray field error /tildewideh. In fact, we have14 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN Lemma 1. For any 1≤r <∞, andn= 2,3, it holds that (56)/vextenddouble/vextenddoubleMεhd[(Mε−Mh)m0]−MεHd[Ms(y)](x ε)·m0/vextenddouble/vextenddouble Lr(Ω) ≤Cε1/r+Cεln(ε−1+1), whereCdepends on /ba∇dbl∇m0/ba∇dblW1,∞(Ω),/ba∇dbl∇Ms(y)/ba∇dblH1(Y)and is independent of ε. Proof of Lemma 1will be given in Section 4.2. Lemma 1directly leads to the consistency error: Theorem 3. (Consistency) Given fεdefined in (25), it can be divided as fε=f0+/tildewidef, satisfying /ba∇dbl/tildewidef/ba∇dblL2(Ω)≤Cε, and (57)/ba∇dblf0/ba∇dblL2(Ω)= 0,whenµ0= 0, /ba∇dblf0/ba∇dblLr(Ω)≤Crµ0/parenleftbig ε1/r+εln(ε−1+1)/parenrightbig ,whenµ0>0,n/ne}ationslash= 1, for any1≤r <+∞. Here constant CandCrdepend on /ba∇dbl∇m0/ba∇dblH4(Ω), /ba∇dbl∇Ms(y)/ba∇dblH1(Y), and are independent of ε. 4.1.Estimate of some singular integral. The strategy to prove Lemma 1is to rewrite the stray field into derivatives of Newtonian po tential, thus the consistency estimate turns into the estimate of singula r integrals. The following Lemmas introduce the estimate of singular integr al in terms of distribution function and boundary layer. We will use the cu t-off function ηεwithin the interior of area away from boundary: (58) 0≤ηε≤1,|∇ηε| ≤Cε−1, ηε(x) = 1 if dist( x,∂Ω)≥2 3ε, ηε(x) = 0 if dist( x,∂Ω)≤1 3ε. where dist( x,∂Ω) denotes the distance between xand∂Ω, and cut-off func- tionφεin a small ball: (59) 0≤φε≤1,|∇φε| ≤Cε−1, φε(x) = 1 if |x| ≤1 3ε, φε(x) = 0 if |x| ≥2 3ε. Denote the boundary layer Ωεas Ωε={x∈Ω,dist(x,∂Ω)≤ε}. Lemma 2. Assume that scalar functions f(y)∈C1(Rn)isY-periodic, g(x)∈C1(¯Ω), define for x∈Ω u(x) =/integraldisplay Ω/vextendsingle/vextendsinglef(x ε)−f(z ε)/vextendsingle/vextendsingle |x−z|ndz, v(x) =/integraldisplay Ωε|g(x)−g(z)| |x−z|ndz, thenu(x)∈L∞(Ω)logarithmically grows with respect to ε, satisfying /ba∇dblu/ba∇dblL∞(Ω)≤Cln(ε−1+1)/ba∇dblf(y)/ba∇dblL∞(Y)+C/ba∇dbl∇f(y)/ba∇dblL∞(Y),HOMOGENIZATION OF THE LLG EQUATION 15 andv(x)∈Lr(Ω)decreases at speed of O(ε1/r)for any1≤r <∞, satisfying /ba∇dblv/ba∇dblLr(Ω)≤Cε1/rln(ε−1+1)/parenleftbig /ba∇dblg(x)/ba∇dblL∞(Ω)+ε/ba∇dbl∇g(x)/ba∇dblL∞(Ω)/parenrightbig . Constant Cis independent of ε. Proof.Splitting the integral in uinto/integraltext Ω−B(x,ε)+/integraltext B(x,ε), one can estimate it by |u(x)| ≤C/integraldisplay Ω−B(x,ε)/ba∇dblf(y)/ba∇dblL∞(Y) |x−z|ndz+Cε−1/integraldisplay B(x,ε)/ba∇dbl∇f(y)/ba∇dblL∞(Y) |x−z|n−1dz, therefore the estimate of uin Lemma follows by simple integral. As for the estimate of v, by application of cut-off function φε=φε(x−z), one has |v(x)|=/integraldisplay Ωεφε|g(x)−g(z)| |x−z|ndz+/integraldisplay Ωε(1−φε)|g(x)−g(z)| |x−z|ndz ≤C/ba∇dbl∇g/ba∇dblL∞(Ω)/integraldisplay Ωεφε |x−z|n−1dz+C/ba∇dblg/ba∇dblL∞(Ω)/integraldisplay Ωε1−φε |x−z|ndz =R1+R2. ForR1, one can write by Fubini’s Theorem /ba∇dblR1/ba∇dblr Lr(Ω)≤C/ba∇dbl∇g/ba∇dblr L∞(Ω)/integraldisplay Ω/parenleftBig/integraldisplay Ωεφε |x−z|n−1dz/parenrightBigr dx ≤C/ba∇dbl∇g/ba∇dblr L∞(Ω)sup x∈Ω/parenleftBig/integraldisplay Ωεφε |x−z|n−1dz/parenrightBigr−1 ×sup z∈Ωε/integraldisplay Ωφε |x−z|n−1dx/integraldisplay Ωε1dz ≤C/ba∇dbl∇g/ba∇dblr L∞(Ω)·Cεr−1·Cε·Cε. As forR2, applying the same argument leads to /ba∇dblR2/ba∇dblr Lr(Ω)≤C/ba∇dblg/ba∇dblr L∞(Ω)sup x∈Ω/parenleftBig/integraldisplay Ωε1−φε |x−z|ndz/parenrightBigr−1 ×sup z∈Ωε/integraldisplay Ω1−φε |x−z|ndx/integraldisplay Ωε1dz ≤C/ba∇dblg/ba∇dblr L∞(Ω)·C[ln(ε−1+1)]r−1·Cln(ε−1+1)·Cε. /square Lemma 3. Assume that a scalar function fε(x)∈L∞(Ω)satisfiesfε(x) = 0whenx∈Ω−Ωε, which means fεis nonzero only in boundary layer. Let w(x)be the Newtonian potential of fεinΩ, i.e., w(x) =/integraldisplay ΩΦ(x−z)fε(z)dz,x∈Ω, whereΦis the fundamental solution of Laplace’s equation. Then w(x)∈ W2,p(Ω)satisfies for any 1≤p <+∞ /ba∇dbl∇2w/ba∇dblLp(Ω)≤C/parenleftbig ε1/p+εln(ε−1+1)/parenrightbig/parenleftbig /ba∇dblfε(x)/ba∇dblL∞(Ω)+ε/ba∇dbl∇fε(x)/ba∇dblL∞(Ω)/parenrightbig .16 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN Constant Cis independent of ε. Proof.The case of 1 < p <+∞follows directly by the property of Newto- nian potential: /ba∇dbl∇2w/ba∇dblLp(Ω)≤C/ba∇dblfε/ba∇dblLp(Ω)≤C/ba∇dblfε/ba∇dblLp(Ωε), and the fact (60) /ba∇dblfε/ba∇dblLp(Ωε)≤ |Ωε|1/p/ba∇dblfε/ba∇dblL∞(Ω). Now let us consider the case of p= 1 and write ∂2w ∂xi∂xj=/integraldisplay Ω∂2 ∂xi∂xj/braceleftbig Φ(x−z)/bracerightbig ·/braceleftbig fε(z)−fε(x)/bracerightbig dz +fε(x)/integraldisplay ∂Ωνi·∂ ∂xj/braceleftbig Φ(x−z)/bracerightbig dz =:S1+S2. ForS1, one can apply Lemma 2to derive /ba∇dblS1/ba∇dblL1(Ω)≤Cεln(ε−1+1)/parenleftbig /ba∇dblfε(x)/ba∇dblL∞(Ω)+ε/ba∇dbl∇fε(x)/ba∇dblL∞(Ω)/parenrightbig . As forS2, we can split the integral into/integraltext ∂Ω−B(x,ε)+/integraltext ∂Ω∩B(x,ε), and write /ba∇dblS2/ba∇dblL1(Ω)≤sup x∈Ω/integraldisplay ∂Ω−B(x,ε)νi·∂ ∂xj/braceleftbig Φ(x−z)/bracerightbig dz×/integraldisplay Ωfε(x)dx + sup z∈∂Ω/integraldisplay Ω∩B(z,ε)νi·∂ ∂xj/braceleftbig Φ(x−z)/bracerightbig ·fε(x)dx×/integraldisplay ∂Ω1dz ≤Cln(ε−1+1)×ε/ba∇dblfε/ba∇dblL∞(Ω)+Cε/ba∇dblfε/ba∇dblL∞(Ω), here in the second line we have used the Fubini’s theorem. Thu s the Lemma is proved. /square 4.2.Consistency error of stray field. Now we are ready to prove the consistency error of stray field /tildewidehin Lemma 1. The idea is to use result in [10] and Green’s representation formula, to rewrite /tildewidehinto singular integral that of the types estimated in above Lemmas. Proof.(Proof of Lemma 1) Recall from ( 2) the stray field in LLG equation can be calculated by (61) hd[(Mε−Mh)m0] =∇U, whereU=U[(Mε−Mh)m0] satisfies ∆U=−div[(Mε−Mh)m0XΩ] inD′(Rn). Denotes the ith component of m0bym0,i. Using the fact |m0|= 1, one can write [10] U(x) =−n/summationdisplay i=1/integraldisplay Ω∂ ∂xiΦ(x−z)(Ms(z ε)−Mh)m0,i(z)dz.HOMOGENIZATION OF THE LLG EQUATION 17 Substituting above representation of U(x) into (61) and making the use of cut-off function ηεdefined in ( 58), one can derive (62)hd[(Mε−Mh)m0] =−∇/parenleftBign/summationdisplay i=1/integraldisplay Ω∂ ∂xiΦ(x−z)ηε(z)(M(z ε)−Mh)m0,i(z)dz/parenrightBig −∇/parenleftBign/summationdisplay i=1/integraldisplay Ωε∂ ∂xiΦ(x−z)(1−ηε(z))(M(z ε)−Mh)m0,i(z)dz/parenrightBig =:Pε+/tildewidePε, where/tildewidePεis the derivative of Newtonian potential in boundary layer t hat can be estimated by Lemma 3. Define/tildewideU(y) as the solution of (63) ∆/tildewideU(y) =−(Ms(y)−Mh), U(y) isY-periodic in y, then one can write from ( 7) and (63) that (64)Hd[Ms(y)](x ε) =ε2∇2/tildewideU(x ε) =ε2∇2/braceleftbig ηε(x)/tildewideU(x ε)/bracerightbig +ε2∇2/braceleftbig (1−ηε(x))/tildewideU(x ε)/bracerightbig . Note that by Green’s representation formula, ε2ηε(x)/tildewideU(x ε) =−/integraldisplay ΩΦ(x−z)∆/parenleftbig ε2ηε(z)/tildewideU(z ε)/parenrightbig dz. Substituting the above formula into ( 64) and using the fact of /tildewideU −∆/parenleftbig ε2ηε(z)/tildewideU(z ε)/parenrightbig =ηε(M(z ε)−Mh)−/braceleftbig ε2∆ηε(z)·/tildewideU(z ε)+2ε2∇ηε(z)·∇/tildewideU(z ε)/bracerightbig , we finally obtain m0·Hd[Ms(y)](x ε) =m0·∇2/integraldisplay ΩΦ(x−z)ηε(z)(M(z ε)−Mh)dz+m0·/braceleftBig ε2∇2/braceleftbig (1−ηε(x))/tildewideU(x ε)/bracerightbig +∇2/integraldisplay ΩΦ(x−z)/braceleftbig ε2∆ηε(z)·/tildewideU(z ε)+2ε2∇ηε(z)·∇/tildewideU(z ε)/bracerightbig dz/bracerightBig =:Qε+/tildewideQε, where the boundary layer term /tildewideQεcan be estimated by Lemma 3and (60). Now in order to estimate the left-hand side of ( 56) in the Lemma, it only remains to consider the term Pε−Qε. Notice that one can write (65)Pε−Qε=n/summationdisplay i=1/integraldisplay Ω∂ ∂xi/braceleftbig ∇xΦ(x−z)/bracerightbig ηε(z) ×(M(z ε)−Mh)/parenleftbig m0,i(x)−m0,i(z)/parenrightbig dz.18 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN With the notation ( 63), one has (M(z ε)−Mh) =∇z·/braceleftbig ∇z(ε2/tildewideU(z ε))−∇x(ε2/tildewideU(x ε))/bracerightbig . After substituting it into ( 65) and applying integration by parts, the leading integrals are estimated directly by application of Lemma 2. /square 5.Boundary Corrector 5.1.Neumann corrector. Let usgive thedefinition of Neumanncorrector ωNas (66) ωN=n/summationdisplay i=1/parenleftbig Φi−xi−εχε i/parenrightbig∂m0 ∂xi with the notation χε i(x) =χi(x ε),xiis theith component of spatial variable, and (Φ i)1≤i≤nis given by (67) div(aε∇Φi) = div(a0∇xi) in Ω, ∂ ∂νεΦi=∂ ∂νhxion∂Ω. Here we denote∂ ∂νε=ν·aε∇,∂ ∂νh=ν·a0∇. Thusxiis the homogenized solution of Φ ifrom (67). Since Φ iis unique up to a constant, one may assume Φ i(˜x)−˜x= 0 for some ˜x∈Ω. We introduce that Φ i−xi−εχε ihas following property. Lemma 4. ForΦigiven in (67), under the smoothness assumption on A(y) and∂Ω, it holds that (see [12]) (68) /ba∇dbl∇Φi−∇xi−ε∇χε i/ba∇dblL∞(Ω)≤C,/ba∇dbl∇2Φi/ba∇dblL∞(Ω)≤C, and (69) /ba∇dblΦi−xi/ba∇dblL∞(Ω)≤Cεln(ε−1+1), whereCis independent of ε. Proof.In fact, one has the estimate /vextendsingle/vextendsingle∇Φi−∇xi−ε∇χε i/vextendsingle/vextendsingle≤Cmax{1,ε[dist(x,∂Ω)]−1} from Lemma 7 .4.5 in [12]. This, together with the fact Φ i(˜x)−˜x= 0 , yields (69) by following integrals: |Φi(x)−xi|=/vextendsingle/vextendsingle/vextendsingle/integraldisplay1 0d ds/braceleftBig Φi/parenleftbig ˜x+s(x−˜x)/parenrightbig −/parenleftbig ˜xi+s(xi−˜xi)/parenrightbig/bracerightBig ds/vextendsingle/vextendsingle/vextendsingle ≤C/integraldisplay1 0max{1,ε(1−s)−1}ds≤Cεln(ε−1+1), for anyx∈Ω.HOMOGENIZATION OF THE LLG EQUATION 19 As for the second inequality in ( 68), we prove by making use of the Neu- mann function for operator Aεfrom [12] Section 7.4, denoted by Nε(x,z), and write from ( 67) that (70) Φi(x) =−n/summationdisplay k=1/integraldisplay ∂Ωνk·a0 kiNε(x,z)dz+1 |∂Ω|/integraldisplay ∂ΩΦi(z)dz. Let us denote the projection of∂ ∂xjalong∂ ∂νεbyPxj, and define P⊥ xj= ∂ ∂νε−Pxj, one can write for z∈∂Ω ∂ ∂zjNε(x,z) =/parenleftbig Pzj+P⊥ zj/parenrightbig Nε(x,z) =P⊥ zjNε(x,z). Now applying∂2 ∂xl∂xjto both sides of ( 70), using above formula and integra- tion by parts on ∂Ω lead to (71)∂2 ∂xl∂xjΦi(x) =−n/summationdisplay k=1/integraldisplay ∂ΩP⊥ zlP⊥ zjνk(z)·a0 kiNε(x,z)dz, here we have used the fact that P⊥ zlis a tangential derivative on the bound- ary, and Nε(x,z) =Nε(z,x) by the symmetry of Aε. (71) implies the second inequality in ( 68) by the smoothness assumption of boundary. /square 5.2.A high-order modification. As noted in Section 1, we use ωNto control the Neumann boundary data, and use a modification fun ctionωM to control the inhomogeneous term that induced by ωN, written in ( 15) separately. In order to explain the construction of the modi fication function, we point out that there are some bad terms appear whenwe calcu lateLεωN, which have no convergence in L2norm. Denote the bad terms by T1 badand T2 bad, then they can be written as (72)T1 bad=2n/summationdisplay i,j,k=1∂ ∂xk/braceleftBig aε ki/parenleftbig Φj−xj−εχε j/parenrightbig ·∂2m0 ∂xi∂xj/bracerightBig −n/summationdisplay i,j,k=1/braceleftBig∂ ∂xkaε ik·/parenleftbig Φj−xj−εχε j/parenrightbig/bracerightBig∂2m0 ∂xi∂xj, and T2 bad=αn/summationdisplay i,j=1/parenleftbig aε ij∂ωN ∂xi·∂{2/tildewidemε+ωN} ∂xj/parenrightbig/parenleftbig/tildewidemε+ωN/parenrightbig . Notice that these terms cannot converge for the existence of∂ωN ∂xiand∂aε ik ∂xk. Now let us rewrite T1 badandT2 badinto divergence form up to a small term. ForT1 bad, notice that/summationtextn k=1∂aε ik ∂xk=Aε(εχε i) from (33), substitute it into the20 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN second term on the right-hand side of ( 72), it leads to (73)T1 bad=n/summationdisplay k,l=1∂ ∂xk/parenleftbig aε klG1 l(x)/parenrightbig +n/summationdisplay i,j=1εχε i·Aε/braceleftBig/parenleftbig Φj−xj−εχε j/parenrightbig ·∂2m0 ∂xi∂xj/bracerightBig , whereG1 l(x) in the divergence term reads G1 l(x) =2n/summationdisplay j=1/parenleftbig Φj−xj−εχε j/parenrightbig∂2m0 ∂xl∂xj+n/summationdisplay i,j=1/braceleftBig∂ ∂xl(εχε i) ·/parenleftbig Φj−xj−εχε j/parenrightbig −εχε i·∂ ∂xl/parenleftbig Φj−xj−εχε j/parenrightbig/bracerightBig∂2m0 ∂xi∂xj. As forT2 bad, a direct calculation implies it can be rewritten as T2 bad=n/summationdisplay k,l=1∂ ∂xk/parenleftbig aε klG2 l(x)/parenrightbig −αn/summationdisplay i,j=1aε ij/parenleftbig ωN·∂{2/tildewidemε+ωN} ∂xj/parenrightbig/parenleftbig /tildewidemε+ωN/parenrightbig −αn/summationdisplay i,j=1/parenleftbig ωN·Aε{2/tildewidemε+ωN}/parenrightbig/parenleftbig /tildewidemε+ωN/parenrightbig , (74) whereG2 l(x) in the divergence term can be calculated by G2 l(x) =α/parenleftbig ωN·∂{2/tildewidemε+ωN} ∂xl/parenrightbig/parenleftbig /tildewidemε+ωN/parenrightbig . Moreover, one can apply Lemma 4to deduce from ( 73) and (74) that for i= 1,2 (75)/vextenddouble/vextenddoubleTi bad−n/summationdisplay k,l=1∂ ∂xk/parenleftbig aε klGi l(x)/parenrightbig/vextenddouble/vextenddouble L2(Ω)≤Cε[ln(ε−1+1)]2, here we have use the fact Aε(Φi−xi−εχε i) = 0. Constant Cdepends on /ba∇dbl∇m0/ba∇dblW2,∞(Ω),/ba∇dblAε/tildewidemε/ba∇dblL∞(Ω), but is independent of ε. Now we define the modification function ωM=ω1 M+ω2 M, whereωi M, i= 1,2 satisfies (76) Aεωi M=n/summationdisplay k,l=1∂ ∂xk/parenleftbig aε klGi l(x)/parenrightbig in Ω, ∂ ∂νεωi M=n/summationdisplay k,l=1νk·aε klGi l(x) on∂Ω, hereνkis thek-th component of vector ν. By the Lax-Milgram theorem, onecan obtain theexistence anduniquenessof ωi M,i= 1,2 upto aconstant. Let/integraltext ∂Ωωi Mdx= 0, then the correctors yield the following estimate.HOMOGENIZATION OF THE LLG EQUATION 21 Lemma 5. Forωi M,i= 1,2defined in (76), under smooth assumption of m0and∂Ω, it holds that for n≤3 (77) /ba∇dblωi M/ba∇dblL∞(Ω)≤Cε,/ba∇dbl∇ωi M/ba∇dblL∞(Ω)≤Cεln(ε−1+1), whereCdepends on /ba∇dbl∇m0/ba∇dblW3,∞(Ω)and is independent of ε. Proof.Here we use the Neumann function Nε(x,z) for operator Aε, see [12] Section 7.4. ( 76) implies for i= 1,2 ωi M=n/summationdisplay k,l=1/integraldisplay Ωaε kl∂ ∂zk/braceleftbig Nε(x,z)/bracerightbig Gi l(z)dz. Using the fact ∇zNε(x,z)≤C|x−z|1−n, see [12] p.159, we can derive /ba∇dblωi M/ba∇dblL∞(Ω)≤C/ba∇dblGi l/ba∇dblL∞(Ω)≤Cεln(ε−1+1). As for the second inequality in ( 77), it follows from [ 12], Lemma 7.4.7: /ba∇dbl∇ωi M/ba∇dblL∞(Ω)≤Cln(ε−1+1)/ba∇dblGi l/ba∇dblL∞(Ω)+Cε/ba∇dbl∇Gi l/ba∇dblL∞(Ω) with the estimate /ba∇dbl∇Gi l/ba∇dblL∞(Ω)≤C, from Lemma 4. Here constant Cdepends on /ba∇dbl∇m0/ba∇dblW3,∞(Ω),/ba∇dbl∇(Φj−xj− εχε j)/ba∇dblL∞(Ω), but is independent of εby Lemma 4. /square 5.3.Estimates of initial-boundary data. Theorem 4. Foreε bgiven in (29), withωb=ωN−ωMgiven in (66), under the smooth assumption, it holds that initial data of eε bsatisfies (78) /ba∇dbleε b(x,0)/ba∇dblH1(Ω)≤Cεln(ε−1+1), whereCdepends on /ba∇dbl∇2m0 init/ba∇dblH1(Ω)and is independent of ε. And for the boundary data, it holds that /ba∇dbl∂ ∂νεeε b/ba∇dblW1,∞(0,T;B−1/2,2(∂Ω))≤Cεln(ε−1+1), (79) whereCdepends on /ba∇dbl∇2m0/ba∇dblW1,∞(0,T;B−1/2,2(∂Ω))and is independent of ε. Proof.We rewrite eε bfrom its definition as (80) eε b=mε−m0−n/summationdisplay i=1(Φi−xi)∂m0 ∂xi−ε2m2+ωM. First, let us prove ( 78). By the initial condition of mεandm0, along with the smoothness condition, one can check (81)eε b(x,0) =mε init−m0 init−n/summationdisplay i=1(Φi−xi)∂m0 init ∂xi−ε2m2,init+ωM,init with notation m2,init,ωM,initdefined the same as m2,ωMexcept we replace m0bym0 init. From the assumption ( 16)-(17),m0 initis the homogenized22 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN solution of mε init, by classical homogenization theorem of elliptic problems in [12], one has /ba∇dblmε init−m0 init−(Φ−x)∇m0 init/ba∇dblH1(Ω)≤Cεln(ε−1+1). Also note that by definition of m2and Lemma 5, one has /ba∇dblε2m2,init/ba∇dblH1(Ω)+/ba∇dblωM,init/ba∇dblH1(Ω)≤Cεln(ε−1+1)/ba∇dbl∇2m0 init/ba∇dblH1(Ω). Therefore the inequality ( 78) follows from ( 81) and above estimates. Notice that by the boundary condition of m0in (10) and Φ kin (67), we have n/summationdisplay k=1∂ ∂νε(Φk−xk)·∂m0 ∂xk=−∂ ∂νεm0,x∈∂Ω, therefore applying∂ ∂νεto both sides of ( 80) leads to (82) ∂ ∂νεeε b=−n/summationdisplay k=1(Φk−xk)·∂ ∂νε(∂m0 ∂xk)−ε2∂ ∂νεm2+∂ ∂νεωM,x∈∂Ω. Under the smoothness assumption of m0andaε, we can also derive the smoothness of ( Φ−x),m2andωMover¯Ω. Thus by Lemma 4and Lemma 5, one can directly obtain from ( 82) /ba∇dbl∂ ∂νεeε b/ba∇dblB−1/2,2(∂Ω)≤C/ba∇dblΦ−x/ba∇dblL∞(Ω)+Cε2/ba∇dbl∇m2/ba∇dblL∞(Ω)+C/ba∇dbl∇ωM/ba∇dblL∞(Ω) ≤Cεln(ε−1+1), whereCdepends on /ba∇dbl∇2m0/ba∇dblB−1/2,2(∂Ω)and/ba∇dbl∇2(∂tm0)/ba∇dblB−1/2,2(∂Ω). The same argument for /ba∇dbl∂ ∂νε(∂teε b)/ba∇dblB−1/2,2(∂Ω)leads to ( 79). /square 5.4.Estimates of inhomogeneous terms. Fromtheabovedefinitionand property, we get the main result of this section. Theorem 5. Foreε bgiven in (29), withωb=ωN−ωMgiven in (66)and Lεdefined in (27), under the smooth assumption, it holds that /ba∇dbl∂tωb/ba∇dblL2(Ω)≤Cεln(ε−1+1), (83) /ba∇dblLεωb/ba∇dblL2(Ω)≤Cε[ln(ε−1+1)]2+C/ba∇dblmε−/tildewidemε−ωb/ba∇dblH1(Ω), (84) whereCdepends on /ba∇dblmε/ba∇dblH1(Ω),/ba∇dbl∇2m0/ba∇dblW2,∞(Ω),/ba∇dbl∇(∂tm0)/ba∇dblW1,∞(Ω)and is independent of ε. Moreover, one has the estimate (85) /ba∇dblAεeε b/ba∇dblL2(Ω)≤Cln(ε−1+1), whereCdepends on /ba∇dblAεmε/ba∇dblL2(Ω),/ba∇dbl∇2m0/ba∇dblW2,∞(Ω),/ba∇dbl∇(∂tm0)/ba∇dblW1,∞(Ω)and is independent of ε.HOMOGENIZATION OF THE LLG EQUATION 23 Proof.Inordertoestimateleft-handsideof ( 84), wesplititas /ba∇dblLεωb/ba∇dblL2(Ω)= /ba∇dblLεωN−LεωM/ba∇dblL2(Ω)≤R1+R2+R3, with R1=/vextenddouble/vextenddoubleLεωN−/braceleftbig AεωN−mε×AεωN−D2(ωN)/bracerightbig/vextenddouble/vextenddouble L2(Ω), R2=/vextenddouble/vextenddouble/braceleftbig Aεω1 M−mε×Aεω1 M−Aεω2 M/bracerightbig −LεωM/vextenddouble/vextenddouble L2(Ω), R3=/vextenddouble/vextenddoubleAε(ωN−ω1 M)−mε×Aε(ωN−ω1 M)−/parenleftbig D2(ωN)−Aεω2 M/parenrightbig/vextenddouble/vextenddouble L2(Ω). One can check by definition of LεthatR1does not have derivative of ωN, andR2onlycontains first-orderderivativeof ωM, thustheycanbeestimated by Lemma 4and Lemma 5as R1+R2≤Cεln(ε−1+1), whereCdepends on /ba∇dblmε/ba∇dblH1(Ω),/ba∇dbl∇2m0/ba∇dblW2,∞(Ω). As forR3, in the view of (75), it can be bounded from above by /ba∇dblAεωN−T1 bad/ba∇dblL2(Ω)+/ba∇dblmε×(AεωN−T1 bad)/ba∇dblL2(Ω)+/ba∇dblD2(ωN)−T2 bad/ba∇dblL2(Ω). In above terms, the first term can be estimated by applying Aε(Φi−xi− εχε i) = 0 to derive that /ba∇dblAεωN− T1 bad/ba∇dblL2(Ω)≤Cεln(ε−1+ 1), with C independent of ε. The same result holds for the second term. Now let us estimate the last term. We assert that (86)/ba∇dblD2(ωN)−T2 bad/ba∇dblL2(Ω)≤Cεln(ε−1+1)+C/ba∇dblmε−/tildewidemε−ωb/ba∇dblH1(Ω), whereCdepends on /ba∇dbl∇2m0/ba∇dblW1,∞(Ω)and/ba∇dblmε/ba∇dblH1(Ω). In fact, we denote the terms in D2(ωN) that contain derivatives of ωNby/tildewideD2(ωN), then it reads /tildewideD2(ωN) =αn/summationdisplay i,j=1/parenleftbig aε ij∂ωN ∂xi·∂mε ∂xj+aε ij∂/tildewidemε ∂xi·∂ωN ∂xj/parenrightbig mε, and one can check the remaining terms satisfy /ba∇dblD2(ωN)−/tildewideD2(ωN)/ba∇dblL2(Ω)≤C(1+/ba∇dbl∇m0/ba∇dbl2 L∞(Ω))/ba∇dblωN/ba∇dblL2(Ω)≤Cεln(ε−1+1). Substituting mε= (mε−/tildewidemε−ωN)+(/tildewidemε+ωN) into/tildewideD2(ωN), one can write /tildewideD2(ωN) =T2 bad+αn/summationdisplay i,j=1/parenleftbig aε ij∂ωN ∂xi·∂{mε−/tildewidemε−ωN} ∂xj/parenrightbig mε +αn/summationdisplay i,j=1/parenleftbig aε ij∂ωN ∂xi·∂{/tildewidemε+ωN} ∂xj/parenrightbig/parenleftbig mε−/tildewidemε−ωN/parenrightbig . Hence it follows that /ba∇dbl/tildewideD2(ωN)−T2 bad/ba∇dblL2(Ω)≤C(1+/ba∇dbl∇m0/ba∇dbl2 L∞(Ω))/ba∇dblmε−/tildewidemε−ωN/ba∇dblH1(Ω). The assertion is proved.24 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN As for (83), one can deduce from Lemma 4and the proof of Lemma 5to obtain /ba∇dbl∂tωN/ba∇dblL∞(Ω)≤Cεln(ε−1+1), /ba∇dbl∂tωi M/ba∇dblL∞(Ω)≤C/ba∇dbl∂tGi l/ba∇dblL∞(Ω)≤Cεln(ε−1+1), whereCdepends on /ba∇dbl∇(∂tm0)/ba∇dblW1,∞(Ω). In order to prove ( 85), we use Lemma4, Lemma 5and definition of /tildewidemε, to deduce the estimates /ba∇dblAε/tildewidemε/ba∇dblL2(Ω)+/ba∇dblAεωb/ba∇dblL2(Ω)≤Cln(ε−1+1), then (85) follows with some constant Cdepending on /ba∇dblAεmε/ba∇dblL2(Ω). There- fore Theorem is proved. /square 6.Stability Analysis In this section, we will discuss the stability of following i nitial-boundary problem, which is motivated by equation ( 26): (87) ∂te−Lε(e) =Fin Ω, ν·aε∇e=gon∂Ω, e(0,x) =hin Ω. The following two inequalities will be used. The first inequa lity is motivated byW1,pestimate for oscillatory elliptic problem. Lemma 6. Assumeu∈H2(Ω),ν·aε∇u=gon∂Ω, withg∈B−1/2,2(∂Ω), then it holds that for n≤3, /ba∇dbl∇u/ba∇dblL6(Ω)≤C/ba∇dblAεu/ba∇dblL2(Ω)+C/ba∇dblg/ba∇dblB−1/2,2(∂Ω), moreover, if g= 0, then one has for n≤3 /ba∇dbl∇u/ba∇dblL6(Ω)≤C/ba∇dblAεu/ba∇dblL2(Ω). Constant Cis independent of ε. Proof.We refer that Lemma 6is a direct corollary of Theorem 6.3.2 in [ 12]. One can find the proof in [ 12] [Pages 144-152]. /square We also introduce Sobolev inequality with small coefficient w henn= 2. Lemma 7. For any function f∈H2(Ω), one has when n= 2 /ba∇dblf/ba∇dblL∞(Ω)≤Cln(ε−1+1)/ba∇dblf/ba∇dblH1(Ω)+ε/ba∇dblAεf/ba∇dblL2(Ω), where constant Cis independent of ε. Proof.Using the Neumann function, see [ 12] Section 7.4, one has f=−/integraldisplay Ω∇zNε(x,z)·aε∇f(z)dz+1 |∂Ω|/integraldisplay ∂Ωfdz=:P1+P2.HOMOGENIZATION OF THE LLG EQUATION 25 Applying cut-off function φε=φε(x−z), the first term yields by integration by parts P1=−/integraldisplay Ω(1−φε)∇zNε(x,z)·aε∇f(z)dz+/integraldisplay ΩφεNε(x,z)·Aεf(z)dz +/integraldisplay Ω∇zφε(x−z)·Nε(x,z)·aε∇f(z)dz ≤Cε/ba∇dblAεf/ba∇dblL2(Ω)+Cln(ε−1+1)/ba∇dbl∇f/ba∇dblL2(Ω), here in the last line we have used the fact ∇zNε(x,z)≤C|x−y|−1and Nε(x,z)≤C{1+ln[|x−z|−1]}forn= 2, see [ 12] page 159. As for P2, one has by trace inequality that P2≤C/ba∇dblf/ba∇dblH1(Ω). The Lemma is proved. /square Now let us give the stability of system ( 87) in terms of h,g,Fin L∞(0,T;L2(Ω)) and L∞(0,T;H1(Ω)) norm, respectively. 6.1.Stability in L∞(0,T;L2(Ω)). Theorem 6. Lete∈L∞(0,T;H2(Ω))be a strong solution to (87). Assume h∈L2(Ω),g∈L∞(0,T;B−1/2,2(∂Ω)), andF=F1+F2satisfies (88) F1∈L2(0,T;Lσ(Ω)),F2∈L2(0,T;L2(Ω)) withσ= 1whenn= 1,2, andσ= 6/5whenn= 3, then it holds that, for any0≤t≤T (89)/ba∇dble/ba∇dbl2 L∞(0,T;L2(Ω))+/ba∇dbl∇e/ba∇dbl2 L2(0,T;L2(Ω)) ≤Cδ/parenleftBig /ba∇dblh/ba∇dbl2 L2(Ω)+/ba∇dblg/ba∇dbl2 L2(0,T;B−1/2,2(∂Ω))+γ(ε)/ba∇dblF1/ba∇dbl2 L2(0,T;Lσ(Ω))/parenrightBig +δ/ba∇dblF2/ba∇dbl2 L2(0,T;L2(Ω))+ε2/ba∇dblAεe/ba∇dbl2 L2(0,T;L2(Ω)), for any small δ >0, where /braceleftBigg γ(ε) = 1, whenn= 1,3, γ(ε) = [ln(ε−1+1)]2,whenn= 2. Cδis a constant depending on /ba∇dbl∇mε/ba∇dblL4(Ω),/ba∇dbl∇/tildewidemε/ba∇dblL4(Ω), but is independent oftandε. Proof.The inner product between ( 87) andeinL2(Ω) leads to (90)1 2d dt/integraldisplay Ω|e|2dx−α/integraldisplay Ω/tildewideHε e(e)·edx =−/integraldisplay ΩD1(e)·edx−/integraldisplay ΩD2(e)·edx−/integraldisplay ΩF·edx. Now let us give the estimates to ( 90) term by term. Integration by parts for the second term on the left-hand side yields −/integraldisplay Ω/tildewideHε e(e)·edx≥n/summationdisplay i,j=1/integraldisplay Ωaε ij∂e ∂xi·∂e ∂xjdx−/integraldisplay ∂Ωg·edx−C/integraldisplay Ω|e|2dx26 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN with the boundary term satisfying /integraldisplay ∂Ωg·edx≤ /ba∇dblg/ba∇dblB−1/2,2(∂Ω)/ba∇dble/ba∇dblB1/2,2(∂Ω)≤C/ba∇dblg/ba∇dblB−1/2,2(∂Ω)/ba∇dble/ba∇dblH1(Ω). By integration by parts and the same argument for the boundar y term, the first term on the right-hand side can be estimated as −/integraldisplay ΩD1(e)·edx≤C/integraldisplay Ω|e|2dx+δC/integraldisplay Ω|∇e|2dx−C/ba∇dblg/ba∇dblB−1/2,2(∂Ω)/ba∇dble/ba∇dblH1(Ω). For the second term on the right-hand side of ( 90), using the estimates /integraldisplay Ω/parenleftbig Bε[e,mε]/parenrightbig mε·edx≤C/ba∇dbl∇mε/ba∇dblL4(Ω)/ba∇dble/ba∇dblL4(Ω)/ba∇dbl∇e/ba∇dblL2(Ω) +C/ba∇dblmε/ba∇dblL4(Ω)/ba∇dble/ba∇dblL4(Ω)/ba∇dble/ba∇dblL2(Ω),/integraldisplay Ωgε l[/tildewidemε]e·edx≤C/ba∇dbl∇/tildewidemε/ba∇dbl2 L4(Ω)/ba∇dble/ba∇dbl2 L4(Ω)+C/ba∇dbl/tildewidemε/ba∇dbl2 L4(Ω)/ba∇dble/ba∇dbl2 L4(Ω), and the same argument can be applied to the other terms, we fina lly obtain by Sobolev inequality −/integraldisplay ΩD2(e)·edx≤C+C/integraldisplay Ω|e|2dx+δC/integraldisplay Ω|∇e|2dx, whereC=C0/parenleftbig 1+/ba∇dbl∇mε/ba∇dbl2 L4(Ω)+/ba∇dbl∇/tildewidemε/ba∇dbl2 L4(Ω)/parenrightbig . For the last term in ( 90), by the assumption ( 88), we apply Sobolev inequality for n= 1,3, and apply Lemma7forn= 2, it follows that −/integraldisplay ΩF1·edx≤ C/ba∇dblF1/ba∇dbl2 L1(Ω)+δ/ba∇dble/ba∇dbl2 H1(Ω), n = 1 C[ln(ε−1+1)]2/ba∇dblF1/ba∇dbl2 L1(Ω) +δ/ba∇dble/ba∇dbl2 H1(Ω)+ε2/ba∇dblAεe/ba∇dbl2 L2,n= 2 C/ba∇dblF1/ba∇dbl2 L6/5(Ω)+δ/ba∇dble/ba∇dbl2 H1(Ω), n = 3 −/integraldisplay ΩF2·edx≤δ∗/ba∇dblF2/ba∇dbl2 L2(Ω)+C/ba∇dble/ba∇dblL2(Ω), with any small δ,δ∗>0. Substituting above estimates, one can derive from (90) that 1 2d dt/integraldisplay Ω|e|2dx+(αamin−2δ)/integraldisplay Ω|∇e|2dx≤C/integraldisplay Ω|e|2dx+C/ba∇dblg/ba∇dbl2 B−1/2,2(∂Ω) +C[ln(ε−1+1)]2/ba∇dblF1/ba∇dbl2 Lσ(Ω)+δ∗/ba∇dblF2/ba∇dbl2 L2(Ω)+ε2/ba∇dblAεe/ba∇dbl2 L2. Then (89) follows directly by taking δsmall enough, and the application of Gr¨ onwall’s inequality. /squareHOMOGENIZATION OF THE LLG EQUATION 27 6.2.Stability in L∞(0,T;H1(Ω)). Theorem 7. Lete∈L∞(0,T;H2(Ω))be a strong solution to (87). Assume h∈H1(Ω),g∈H1(0,T;B−1/2,2(∂Ω)), andF∈L2(0,T;L2(Ω)), it holds (91) /ba∇dbl∇e/ba∇dbl2 L∞(0,T;L2(Ω))≤C/parenleftbig /ba∇dblh/ba∇dbl2 H1(Ω)+/ba∇dblg/ba∇dbl2 H1(0,T;B−1/2,2(∂Ω))+/ba∇dblF/ba∇dbl2 L2(0,T;L2(Ω))/parenrightbig , whereCdepends on /ba∇dbl∇mε/ba∇dblL4(Ω),/ba∇dbl∇/tildewidemε/ba∇dblL4(Ω)and/ba∇dblHε e(/tildewidemε)/ba∇dblL4(Ω), but is independent of tandε. Proof.The inner product between ( 87) and/tildewideHε e(e) inL2(Ω) leads to (92)−/integraldisplay Ω∂te·/tildewideHε e(e)dx+α/integraldisplay Ω/tildewideHε e(e)·/tildewideHε e(e)dx =/integraldisplay ΩD1(e)·/tildewideHε e(e)dx+/integraldisplay ΩD2(e)·/tildewideHε e(e)dx+/integraldisplay ΩF·/tildewideHε e(e)dx. In the following we give the estimates to ( 92) term by term. Note that integration by parts yields −/integraldisplay Ω∂te·/tildewideHε e(e)dx=d dtGε L[e]−/integraldisplay ∂Ω∂te·gdx, =d dtGε L[e]−∂t/integraldisplay ∂Ωe·gdx−C/ba∇dbl∂tg/ba∇dblB−1/2,2(∂Ω)/ba∇dble/ba∇dblH1(Ω). Using the fact mε×/tildewideHε e(e)·/tildewideHε e(e) = 0, the first term on the right-hand side of (92) can be estimate by Sobolev inequality as /integraldisplay ΩD1(e)·/tildewideHε e(e)dx≤C/ba∇dblHε e(/tildewidemε)/ba∇dblL4(Ω)/ba∇dble/ba∇dblL4(Ω)/ba∇dbl/tildewideHε e(e)/ba∇dblL2(Ω) ≤C/ba∇dble/ba∇dbl2 H1(Ω)+δC/ba∇dbl/tildewideHε e(e)/ba∇dbl2 L2(Ω), whereC=C0/parenleftbig 1+/ba∇dblHε e(/tildewidemε)/ba∇dbl2 L4(Ω)/parenrightbig . For the second term on the right-hand side of (92), note that we have the estimate /integraldisplay Ω/parenleftbig Bε[e,mε]/parenrightbig mε·/tildewideHε e(e)dx≤C/ba∇dbl∇mε/ba∇dblL4(Ω)/ba∇dbl∇e/ba∇dblL4(Ω)/ba∇dbl/tildewideHε e(e)/ba∇dblL2(Ω) +C/ba∇dblmε/ba∇dblL4(Ω)/ba∇dble/ba∇dblL4(Ω)/ba∇dbl/tildewideHε e(e)/ba∇dblL2(Ω), in which one can deduce /ba∇dbl∇e/ba∇dblL4(Ω)≤/ba∇dbl∇e/ba∇dbl1/3 L2(Ω)/ba∇dbl∇e/ba∇dbl2/3 L6(Ω) ≤C/ba∇dbl∇e/ba∇dbl1/3 L2(Ω)(1+/ba∇dbl/tildewideHε e(e)/ba∇dblL2(Ω)+/ba∇dblg/ba∇dblB−1/2,2(∂Ω))2/3, using interpolation inequality and Lemma 6. The other terms can be esti- mated in the same fashion. After the application of Young’s i nequality, one finally obtains /integraldisplay ΩD2(e)·/tildewideHε e(e)dx≤C/ba∇dble/ba∇dbl2 H1(Ω)+δC/ba∇dbl/tildewideHε e(e)/ba∇dbl2 L2(Ω)+C/ba∇dblg/ba∇dbl2 B−1/2,2(∂Ω),28 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN whereC=C0(1+/ba∇dbl∇mε/ba∇dbl2 L4(Ω)+/ba∇dbl∇/tildewidemε/ba∇dbl2 L4(Ω)). Substitutingaboveestimates into (92), we arrive at d dtGε L[e]+(α−Cδ)/integraldisplay Ω|/tildewideHε e(e)|2dx−∂t/integraldisplay ∂Ωe·gdx ≤C/parenleftbig /ba∇dble/ba∇dbl2 H1(Ω)+/ba∇dblF/ba∇dbl2 L2(Ω)+/ba∇dblg/ba∇dblB−1/2,2(∂Ω)+/ba∇dbl∂tg/ba∇dblB−1/2,2(∂Ω)/parenrightbig . Integrating the above inequality over [0 ,t] with 0 < t < T and using the facts/integraldisplay ∂Ωe·gdx≤δ/ba∇dble/ba∇dbl2 H1(Ω)+C/ba∇dblg/ba∇dbl2 B−1/2,2(∂Ω), Gε L[e]≥amin 2/ba∇dbl∇e/ba∇dbl2 L2(Ω)−C, one can finally derive (amin 2−δ)/ba∇dbl∇e(t)/ba∇dbl2 L2(Ω)+(α−Cδ)/integraldisplayt 0/ba∇dbl/tildewideHε e(e)/ba∇dbl2 L2(Ω)dτ ≤C/integraldisplayt 0/parenleftbig /ba∇dble/ba∇dbl2 L2(Ω)+/ba∇dbl∇e/ba∇dbl2 L2(Ω)+/ba∇dblF/ba∇dbl2 L2(Ω)+/ba∇dbl∂tg/ba∇dblB−1/2,2(∂Ω)/parenrightbig dτ+J(h), whereJ(h) yields J(h) =Gε L[h]−/integraldisplay ∂Ωh·∂ ∂νεhdx≤C/ba∇dblh/ba∇dbl2 H1(Ω)+/ba∇dbl∂ ∂νεh/ba∇dbl2 B−1/2,2(∂Ω). (91) is then derived after taking δsmall enough and the application of Gr¨ onwall’s inequality. /square 7.Regularity In the estimate of boundary corrector and stability analysi s by Theorem 5, Theorem 6, Theorem 7, the constant we deduced rely on the value of /ba∇dblAεmε/ba∇dblL2(Ω)and/ba∇dbl∇mε/ba∇dblL6(Ω). In this section, we introduce the uniform regularity on mε, over a time interval independent of ε. For this purpose, we intend to derive a structure-preserving energy inequali ty, in which the degenerate term are kept in the energy. First, let us introduce an interpolation inequality of the e ffective field Hε e(mε) for some S2-valued function mε, which is the generalization of ( 14). The following estimates will be used: a−1 max/ba∇dblmε·Aεmε/ba∇dbl3 L3(Ω)≤ /ba∇dbl∇mε/ba∇dbl6 L6(Ω)≤a−1 min/ba∇dblAεmε/ba∇dbl3 L3(Ω), (93) /ba∇dblAεmε/ba∇dblLp(Ω)−Cp≤ /ba∇dblHε e(mε)/ba∇dblLp(Ω)≤ /ba∇dblAεmε/ba∇dblLp(Ω)+Cp, (94) with 1< p <+∞, here the first line follows from the fact −aε|∇mε|2= mε·Aεmεby|mε|= 1 and assumption of aεin (1), and in second line the estimate ( 4) is used. We also introduce a orthogonal decomposition to an y vectoraas (95) a= (mε·a)mε−mε×(mε×a).HOMOGENIZATION OF THE LLG EQUATION 29 Lemma 8. Givenmε∈H3(Ω)that satisfies |mε|= 1and Neumann bound- ary condition ν·aε∇mε= 0, then it holds for n≤3and any 0< δ <1 (96)/ba∇dblHε e(mε)/ba∇dbl3 L3(Ω)≤Cδ+Cδ/ba∇dblHε e(mε)/ba∇dbl6 L2(Ω)+δ/ba∇dblmε×∇Hε e(mε)/ba∇dbl2 L2(Ω), whereCδis a constant depending on δbut independent of ε. Proof.Applying decomposition ( 95) by taking a=Hε e(mε), one can write (97)/ba∇dblHε e(mε)/ba∇dbl3 L3(Ω)≤/integraldisplay Ω|mε·Hε e(mε)|3dx +/integraldisplay Ω|mε×Hε e(mε)|3dx=:I1+I2. Now let us estimate the right-hand side of ( 97) separately. For I1, we apply (93) and Remark 6to derive I1≤C+C/ba∇dbl∇mε/ba∇dbl6 L6(Ω)≤C+C/ba∇dblHε e(mε)/ba∇dbl6 L2(Ω). As forI2, we have by Sobolev inequality for n≤3 I2≤C+C/ba∇dblmε×Hε e(mε)/ba∇dbl6 L2(Ω)+δ∗/ba∇dblmε×Hε e(mε)/ba∇dbl2 H1(Ω), here in the last term, we can apply ( 93)-(94) to derive: δ∗/ba∇dbl∇mε×Hε e(mε)/ba∇dbl2 L2(Ω)≤δ∗/ba∇dbl∇mε/ba∇dbl6 L6(Ω)+δ∗/ba∇dblHε e(mε)/ba∇dbl3 L3(Ω) ≤C+Cδ∗/ba∇dblHε e(mε)/ba∇dbl3 L3(Ω). Now let us turn back to ( 97), we finally obtain (1−Cδ∗)/ba∇dblHε e(mε)/ba∇dbl3 L3(Ω)≤C+C/ba∇dblHε e(mε)/ba∇dbl6 L2(Ω)+δ∗/ba∇dblmε×∇Hε e(mε)/ba∇dbl2 L2(Ω). Letδ∗<1 2C, one can derive ( 96) withδ=δ∗/(1−Cδ∗)<1. /square Now let us recall some energy property of LLG equation, and gi ve the uniform regularity result. Using the formula of vector oute r production (98) a×(b×c) = (a·c)b−(a·b)c, one can rewrite LLG equation ( 22) into a degenerate form (99) ∂tmε+αmε×/parenleftbig mε×Hε e(mε)/parenrightbig +mε×Hε e(mε) = 0. Multiplying ( 99) byHε e(mε) and integrating over (0 ,t), we derive the energy dissipation identity (100) Gε L[mε(t)]+α/integraldisplayt 0/ba∇dblmε×Hε e(mε)/ba∇dbl2 L2(Ω)dτ=Gε L[mε(0)], together ( 99) and (100) leads to the integrable of kinetic energy α 1+α2/integraldisplayt 0/ba∇dbl∂tmε/ba∇dbl2 L2(Ω)dτ≤ Gε L[mε(0)]. (101) Theenergyidentity ( 100)impliestheuniformregularityof /ba∇dblmε×Aεmε/ba∇dbl2 L2(Ω), however, this is not enough to obtain the regularity of /ba∇dblAεmε/ba∇dbl2 L2(Ω)due to the degeneracy. In this end we introduce that:30 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN Theorem 8. Letmε∈L2([0,T];H3(Ω))be a solution to (8). Assume n≤ 3, then there exists T∗∈(0,T]independent of ε, such that for 0≤t≤T∗, /ba∇dblAεmε(t)/ba∇dbl2 L2(Ω)+/integraldisplayt 0/vextenddouble/vextenddoublemε×∇Hε e(mε)(τ)/vextenddouble/vextenddouble2 L2(Ω)dτ≤C, and therefore, by the Sobolev-type inequality in Remark 6, /ba∇dbl∇mε(·,t)/ba∇dbl2 L6(Ω)≤C, whereCis a constant independent of εandt. Proof.Applying ∇to (99) and multiplying by aε∇Hε e(mε) lead to (102)−/integraldisplay Ω∇(∂tmε)·aε∇Hε e(mε)dx =α/integraldisplay Ω∇/parenleftbig mε×(mε×Hε e(mε))/parenrightbig ·aε∇Hε e(mε)dx +n/summationdisplay i,j=1/integraldisplay Ω∂ ∂ximε×Hε e(mε)·aε ij∂ ∂xjHε e(mε)dx=:J1+J2. Denote Γε(mε) =Hε e(mε)−Aεmε. After integration by parts, the left-hand side of (102) becomes −/integraldisplay Ω∇(∂tmε)·aε∇Hε e(mε)dx=/integraldisplay ΩAε(∂tmε)·/parenleftbig Aεmε+Γε(mε)/parenrightbig dx, where the right-hand side can be rewritten as 1 2d dt/integraldisplay Ω|Aεmε|2dx+d dt/integraldisplay ΩAεmε·Γε(mε)dx−/integraldisplay ΩAε(mε)·Γε(∂tmε)dx. Now let us consider the right-hand side of ( 102). ForJ1, one can derive by swapping the order of mixed product J1=−αn/summationdisplay i,j=1/integraldisplay Ω/parenleftbig mε×∂ ∂xiHε e(mε)/parenrightbig ·aε ij/parenleftbig mε×∂ ∂xjHε e(mε)/parenrightbig dx+F1, herethefirstterm onright-handsideissign-preservedduet otheuniformco- erciveness of aεin (1). As for J2, we apply ( 95) by taking a=aε ij∂jHε e(mε), it leads to (103)J2=n/summationdisplay i,j=1/integraldisplay Ωmε×/parenleftbig∂ ∂ximε×Hε e(mε)/parenrightbig ·/parenleftbig mε×aε ij∂ ∂xjHε e(mε)/parenrightbig dx −n/summationdisplay i,j=1/integraldisplay Ω/parenleftbig mε×Hε e(mε)·∂ ∂ximε/parenrightbig mε·aε ij∂ ∂xjHε e(mε)dx.HOMOGENIZATION OF THE LLG EQUATION 31 Usingpropertyof vector outer production( 98) forfirstterm, andintegration by parts for the second term, ( 103) becomes J2=2n/summationdisplay i,j=1/integraldisplay Ω/parenleftbig mε·Hε e(mε)/parenrightbig/parenleftbig mε×aε ij∂ ∂xjHε e(mε)·∂ ∂ximε/parenrightbig dx+F2 ≤C/ba∇dbl∇mε/ba∇dbl6 L6(Ω)+C/ba∇dblHε e(mε)/ba∇dbl3 L3(Ω)+δ/ba∇dblmε×∇Hε e(mε)/ba∇dbl2 L2(Ω)+F2. Here low-order terms Fi,i= 1,2 satisfies by ( 94) and H¨ older’s inequality Fi≤C+C/ba∇dbl∇mε/ba∇dbl6 L6(Ω)+C/ba∇dblHε e(mε)/ba∇dbl3 L3(Ω). Substituting above estimates into ( 102), applying estimate ( 93) and Lemma 8, we finally arrive at (104)1 2d dt/ba∇dblAεmε/ba∇dbl2 L2(Ω)+(αamin−Cδ)/ba∇dblmε×∇Hε e(mε)/ba∇dbl2 L2(Ω) ≤C+C/ba∇dblAεmε/ba∇dbl6 L2(Ω)+C/ba∇dbl∂tmε/ba∇dbl2 L2(Ω)−d dt/integraldisplay ΩAεmε·Γε(mε)dx, Integrating ( 104) over [0 ,t], using the integrability of kinetic energy ( 101) and the following inequality /integraldisplay ΩAεmε·Γε(mε)dx≤C/ba∇dblΓε(mε)/ba∇dbl2 L2(Ω)+1 4/ba∇dblAεmε/ba∇dbl2 L2(Ω), one has for any t∈(0,T] (105)1 4/ba∇dblAεmε(t)/ba∇dbl2 L2(Ω)≤C+C/integraldisplayt 0/ba∇dblAεmε(τ)/ba∇dbl6 L2(Ω)dτ, whereCdependson /ba∇dblAεmε init/ba∇dblL2(Ω),Gε L[mε init] thus is independent of εandt by assumption ( 16)-(17) and Lemma 6. Denote the right-hand side of ( 105) byF(t) and write d dtF(t)≤CF3(t). By the Cauchy-Lipshitz-Picard Theorem [ 3] and comparison principle, there existsT∗∈(0,T] independent of ε, such that F(t) is uniformly bounded on [0,T∗], thus/ba∇dblAεmε(t)/ba∇dbl2 L2(Ω)is uniformly bounded by ( 105). The Lemma is proved. /square Acknowledgments J. Chen was supported by National Natural Science Foundatio n of China via grant 11971021. J.-G. Liu was supported by Natural Scien ce Foun- dation via grant DMS-2106988. Z. Sun was supported by the Pos tgradu- ate Research & Practice Innovation Program of Jiangsu Provi nce via grant KYCX21 2934.32 JINGRUN CHEN, JIAN-GUO LIU, AND ZHIWEI SUN References 1. Francois Alouges, AnneDeBouard, Benoˆ ıt Merlet, andL´ e aNicolas, Stochastic homog- enization of the Landau-Lifshitz-Gilbert equation , Stochastics and Partial Differential Equations: Analysis and Computations 9(2021), 789–818. 2. Francois Alouges and Giovanni Di Fratta, Homogenization of composite ferromagnetic materials , Proceedings of the Royal Society A: Mathematical, Physica l and Engineer- ing Sciences 471(2015), 20150365. 3. Haim Brezis and Haim Br´ ezis, Functional analysis, sobolev spaces and partial differ- ential equations , vol. 2, Springer, 2011. 4. Jingrun Chen, Rui Du, Zetao Ma, Zhiwei Sun, and Zhang Lei, On the multiscale Landau-Lifshitz-Gilbert equation: Two-scale convergenc e and stability analysis , Mul- tiscale Modeling & Simulation (2022), in press. 5. Catherine Choquet, Mohammed Moumni, and Mouhcine Tiliou a,Homogenization of the Landau-Lifshitz-Gilbert equation in a contrasted comp osite medium , Discrete & Continuous Dynamical Systems-S 11(2018), 35. 6. T. L.Gilbert, A Lagrangian formulation of gyromagnetic equation of the ma gnetization field, Physical Review D 100(1955), 1243–1255. 7. Lev Davidovich Landau and E Lifshitz, On the theory of the dispersion of mag- netic permeability in ferromagnetic bodies , Physikalische Zeitschrift der Sowjetunion 8(1935), 153–169. 8. Lena Leitenmaier and Olof Runborg, On homogenization of the Landau-Lifshitz equa- tion with rapidly oscillating material coefficient , Communications in Mathematical Sciences 20(2022), 653–694. 9. ,Upscaling errors in heterogeneous multiscale methods for t he Landau-Lifshitz equation, Multiscale Modeling & Simulation 20(2022), 1–35. 10. Dirk Praetorius, Analysis of the operator delta div arising in magnetic model s, Zeitschrift Fur Analysis Und Ihre Anwendungen 23(2004), 589–605. 11. K´ evin Santugini-Repiquet, Homogenization of the demagnetization field operator in periodically perforated domains , Journal of Mathematical Analysis and Applications 334(2007), 502–516. 12. Zhongwei Shen, Periodic homogenization of elliptic systems , Springer, 2018. School of MathematicalSciences, Universityof Science and Technology of China, Hefei, Anhui 230026, China; Suzhou Institute for Adv anced Research, Universityof Science and Technology of China, Suzhou, Jian gsu 215123, China Email address :jingrunchen@ustc.edu.cn Department of Mathematics and Department of Physics, Duke U niversity, Box 90320, Durham NC 27708, USA Email address :jliu@phy.duke.edu School of Mathematical Sciences, Soochow University, Suzh ou, Jiangsu 215006, China Email address :20194007008@stu.suda.edu.cn | 2022-06-22 | The full Landau-Lifshitz-Gilbert equation with periodic material coefficients
and natural boundary condition is employed to model the magnetization dynamics
in composite ferromagnets. In this work, we establish the convergence between
the homogenized solution and the original solution via a Lax equivalence
theorem kind of argument. There are a few technical difficulties, including: 1)
it is proven the classic choice of corrector to homogenization cannot provide
the convergence result in the $H^1$ norm; 2) a boundary layer is induced due to
the natural boundary condition; 3) the presence of stray field give rise to a
multiscale potential problem. To keep the convergence rates near the boundary,
we introduce the Neumann corrector with a high-order modification. Estimates on
singular integral for disturbed functions and boundary layer are deduced, to
conduct consistency analysis of stray field. Furthermore, inspired by length
conservation of magnetization, we choose proper correctors in specific
geometric space. These, together with a uniform $W^{1,6}$ estimate on original
solution, provide the convergence rates in the $H^1$ sense. | Homogenization of the Landau-Lifshitz-Gilbert equation with natural boundary condition | 2206.10948v1 |
arXiv:1008.0674v1 [cond-mat.mtrl-sci] 3 Aug 2010Determinationofthe spin-flip timeinferromagnetic SrRuO 3from time-resolved Kerr measurements C.L.S.Kantner,1,2M.C.Langner,1,2W.Siemons,3J.L.Blok,4G.Koster,4A.J.H.M.Rijnders,4R.Ramesh,1,3andJ.Orenstein1,2 1Department of Physics, University of California, Berkeley , CA 94720 2Materials Science Division, Lawrence Berkeley National La boratory, Berkeley, CA 94720 3Department of Materials Science and Engineering, Universi ty of California, Berkeley, CA 94720 4MESA+Institute for Nanotechnology, University of Twente, 7500 A E Enschede, The Netherlands (Dated: December 6, 2018) Wereport time-resolvedKerr effectmeasurements of magnet izationdynamics inferromagnetic SrRuO 3. We observe that the demagnetization timeslows substantially at temperatures within15K of theCurie temperature, whichis∼150K. We analyze the data witha phenomenological model that relates the demagnetization timeto the spinfliptime. Inagreement withour observations the mod el yields a demagnetization timethat is inversely proportional toT-T c. Wealsomake adirectcomparisonofthespinfliprateandtheG ilbertdampingcoefficient showing thattheir ratioveryclose tok BTc,indicating a common originfor these phenomena. I: Introduction Thereisincreasinginterestincontrollingmagnetisminfe r- romagnets. Of particular interest are the related question s of howquicklyandbywhatmechanismthemagnetizationcanbe changed by external perturbations. In addition to advancin g our basic understanding of magnetism, exploring the speed withwhichthemagneticstatecanbechangediscrucialtoap- plications such as ultrafast laser-writing techniques. De spite its relevance, the time scale and mechanisms underlying de- magnetizationarenotwell understoodata microscopicleve l. BeforeBeaurepaireetal.’spioneeringworkonlaser-excit ed Ni in 1996,it was thoughtthat spins wouldtake nanoseconds torotate,withdemagnetizationresultingfromtheweakint er- actionofspinswiththelattice. TheexperimentsonNishowe d that this was not the case and that demagnetizationcould oc- cur on time scales significantly less than 1 ps1. Since then demagnetization is usually attributed to Elliott-Yafet me cha- nism, in which the rate of electron spin flips is proportional to the momentum scattering rate. Recently Koopmans et al. have demonstrated that electron-phononor electron-impur ity scattering can be responsible for the wide range of demag- netization time scales observed in different materials2. Also recentlyit hasbeenproposedthat electron-electronscatt ering should be included as well as a source of Elliott-Yafet spin flipping,andconsequently,demagnetization3. AlthoughRef.3 specifically refers to interband scattering at high energie s, it is plausible that intrabandelectron scattering can lead to spin memorylossaswell. Time-resolved magneto-optical Kerr effect (TRMOKE) measurementshavebeendemonstratedtobeausefulprobeof ultrafast laser-induceddemagnetization1. In this paper we re- portTRMOKEmeasurementsonthinfilmsofSRO/STO(111) between 5 and 165K. Below about 80 K we observe damped ferromagneticresonance (FMR), from which we determine a Gilbert damping parameter consistent with earlier measure - ments on SrTiO 3with (001) orientation6. As the the Curie temperature ( ∼150K) is approached the demagnetization time slows significantly, as has been observed in other mag- netic systems4. The slowing dynamics have been attributed to critical slowing down, due to the similarities between th e temperature dependencies of the demagnetization time andthe relaxation time5. In this paper we develop an analytical expression relating the demagnetization time to the spin-fl ip timenearthe Curietemperature. Thisprovidesa newmethod of measuring the spin-flip time, which is essential to under- standingthedynamicsoflaser-induceddemagnetization. II: SampleGrowthandCharacterization SRO thin films were grown via pulsed laser deposition at 700◦C in 0.3 mbar of oxygenand argon(1:1) on TiO 2termi- nated STO(111)7. A pressed pellet of SRO was used for the targetmaterial and the energyon the targetwas kept constan t at 2.1 J/cm2. High-pressure reflection high-energy electron diffraction (RHEED) was used to monitor the growth speed and crystallinity of the SRO film in situ. RHEED patterns andatomicforcemicroscopyimagingconfirmedthepresence of smooth surfaces consisting of atomically flat terraces se p- arated by a single unit cell step (2.2 ˚Ain the [111] direction). X-ray diffraction indicated fully epitaxial films and x-ray re- flectometry was used to verify film thickness. Bulk magneti- zationmeasurementsusingaSQUIDmagnetometerindicated a Curie temperature,T c, of∼155K.Electrical transportmea- surementswere performedin the Vander Pauwconfiguration andshowtheresidualresistanceratiotobeabout10forthes e films. III: ExperimentalMethods IntheTRMOKEtechniqueamagneticsampleisexcitedby theabsorptionofapumpbeam,resultinginachangeofpolar- izationangle, ∆ΘK(t),ofatimedelayedprobebeam. Theul- trashortpulsesfroma Ti:Sapphlaser are used to achievesub - picosecondtime resolution. Near normalincidence,as in th is experiment, ∆ΘKis proportional to the ˆzcomponent of the perturbedmagnetization, ∆Mz.∆ΘKis measured via a bal- anceddetectionscheme. Foradditionalsensitivity,thede riva- tiveof∆ΘKt)withrespecttotimeismeasuredbylockinginto thefrequencyofasmallamplitude( ∼500fs)fastscanningde- lay line in the probe beam path as time is stepped throughon anotherdelayline. IV.1: ExperimentalResults: Low Temperature Fig. 1 shows the time derivative of ∆ΘKfor an 18.5nm SRO/STO(111)sample forthe 16psfollowingexcitationbya pump beam, for temperatures between 5 and 85K. Clear fer- romagnetic resonance (FMR) oscillations are present, gene r-2 FIG.1. DerivativeofthechangeinKerrrotationasafunctio noftime delay followingpulsed photoexcitation, for 5 <T<85 K ated bya suddenshift in easy axisdirectionuponthermal ex- citation by a pump beam6. This motion is described by the Landau-Lifshitz-Gilbertequationwith thefrequencyofos cil- lation proportional to the strength of the magnetocrystall ine anisotropy field, and the damping described by dimension- less phenomenological parameter, α. The motion appears as a decaying oscillation to TRMOKE. The orientation of the anisotropyfield, closer to in-planewith the sample surface in SRO/STO(111) than in SRO/STO(001), makes these oscilla- tions more prominent when observed with the polar Kerr ge- ometrycomparedtopreviousmeasurements. Attempting to model the time derivative of ∆ΘKwith a dampedcosine revealsthat it cannot be fit by such a function for t<2ps. The feature at short times in Fig. 1 contains higherfrequencycomponents,whereastheoscillationswhi ch become clear after 2 ps are at a single frequency. A com- parison of the amplitude of the first peak (at t ∼.5 ps) with the amplitude of the subsequent oscillations (defined as the difference between d ∆ΘK/dt at the peak at ∼3.5 ps and the dip at∼5.5 ps), is shown as a functionof temperaturein Fig. 2. The constant offset between the two amplitudes indicates that d∆ΘK/dt is comprised of a superposition of a tempera- tureindependent,short-livedcomponentwiththelongerli ved dampedoscillations. Fitting the oscillatory portion of the signal to a damped cosine, the temperature dependencies of the amplitude, fre - quency,anddampingparameterarefound,asshowninFig. 3. ComparingtheseparametersforSRO/STO(111)topreviously published work on SRO/STO(001), the frequencyis found to be somewhat smaller and to change more with temperature. Of particular interest is α, which is also smaller in this ori- entation of SRO, consistent with the more pronounced FMR oscillations. Strikingly, in both orientations there is a d ip in αaround45K,whichisrelativelystrongerinSRO/STO(111). This further strengthens the link between αand the anoma- lous hall conductivity, speculated in that paper, through n ear degeneraciesin thebandstructure6. IV.2: ExperimentalResults: HighTemeperatureFIG. 2. Comparing amplitudes of the short time feature and th e fer- romagnetic resonance oscillations By taking the time derivative of ∆ΘK, the FMR oscilla- tionscanbe followeduntilthey disappearat elevatedtempe r- atures,atwhichpointitbecomessimplertolookat ∆ΘKthan its time derivative. Fig. 4 shows ∆ΘKas a function of time for the first 38 ps after excitation by the pump laser, for tem- peratures between 120K and 165K. A property of a second order phase transitions is that the derivative of the order p a- rameter divergesnear the transition temperature. The peak in magnitudeof ∆ΘKin figure 4, shown in figure5, can be un- derstood as the result of the derivative of magnetization wi th respect to temperature becoming steeper near the Curie tem- perature. A strongtemperaturedependenceofthedemagneti - zationtime, τM, is seen,with τMsignificantlyenhancednear 150K,consistentwithpreviousreportsonSRO4,6. ∆ΘK(t) in Fig. 4, normalized by the largest value of ∆ΘK(t) in the first 38 ps, can be fit with the following func- tion: fort <0∆ΘK(t) ∆Θmax(t)= 0 fort >0∆ΘK(t) ∆Θmax(t)=C−Ae−t/τM(1) where the decay time is τM. The resulting τMis plotted as a functionof temperaturein Fig. 6. Notably, τMincreases by a factor of 10 from 135K to 150K. Taking the fit value of Tc= 148.8K, as will be discussed later, τMis plotted log- log as function of reduced temperature, tR= (Tc−T)/Tc. Theresult looksapproximatelylinear,indicatinga powerl aw dependenceof τMonthereducedtemperature. V: Discussion ofResults: Efforts to explain demagnetization have been largely phe- nomenological thus far, understandably, given the dauntin g challenge of a full microscopic model. Beaurepaire et al. in - troduced the three temperature model (3TM) to describe de- magnetization resulting from the interactions of the elect ron, phonon,andspinbaths1. In3TMthedynamicsaredetermined3 FIG. 3. Temperature dependence of (a) Amplitude of oscillat ions, (b) FMRfrequency, and, (c)damping parameter FIG.4. ChangeinKerrrotationasafunctionoftimedelayfol lowing pulsed photoexcitation, for 120 <T<165 KFIG. 5. Magnitude of change in Kerr rotation at 38ps as a funct ion of temperature FIG.6. Demagnetization timeat hightemperature by the specific heats of each bath as well as the coupling constants between them. Demagnetization can generally be described with the appropriate choice of coupling constant s, providing a guide into the microscopic mechanism. Koop- mans et al. also offer a phenomenological description of de- magnetization considering three baths, but one that follow s spin in additionto heat8. Spinis treated asa two state system with energylevels separatedby an exchangegapand Fermi’s goldenrule is usedto relate demagnetizationto electronsc at- teringwhichflipsaspin. Equationsforcouplingconstantsa re derived based on parameters such as the density of states of electrons, phonons, and spins, the electron-phononscatte ring rate,andthe probabilityofspinflipat a scatteringevent. In the following we attempt to understand the behavior of the demagnetization time near T cwith an approach based on the two spin state model. A general relationship between the laser-induced τMand the spin flip time, τsf, can be derived near the transition temperature based on the concept of de- tailed balance9. In equilibrium,the ratio of the probabilityof4 FIG. 7. Log-log plot of demagnetization time as a function of re- duced temperature aspinflippingfrommajoritytominoritytothereverseofthi s process is the Boltzmann factor, e−∆ex/kT, where∆exis the exchange energy gap. The time derivative of the number of majorityandminorityelectronscanthenbewritten: ˙Nmaj=−˙Nmin=Nmin τsf−Nmaj τsfe−∆ex/kBT(2) Whenthesampleisthermallyexcitedbyapumpbeam,the electron temperature is increased by δTe. The rate of change ofspinsisthenalteredinthefollowingway: ˙Nmaj=−˙Nmin=Nmin τsf−Nmaj τsfe−∆ex/kB(T+δTe)(3) The demagnetization time is related to the total change in spin,∆S,frominitialtofinaltemperature,where,setting /planckover2pi1=1, Sis definedby: S= 1/2(Nmaj−Nmin)/Ntotal (4) Assumingthat ∆S,asa functionoftime,canbewritten: ∆S(t) = [S(Tf)−S(Ti)](1−e−t/τM)(5) thedemagnetizationtimecanbewrittenas: τM=∆S ˙S(0)(6) where˙S(0)is the initial change in the time derivative of the spin. The total change in spin can be calculated by taking the derivative of Swith respect to T, and multiplying by ∆Teq, theincreaseintemperatureonceelectrons,phonons,andsp inshavecomeintothermalequilibriumwitheachother. S(T)and ∆Scanbe written: S(T) =−1 2tanh/parenleftbigg∆ 2kT/parenrightbigg (7) and: ∆S=dS dT/vextendsingle/vextendsingle/vextendsingle T=T0∆Teq=−∆ex 4kBT2 0/bracketleftbigg T0∆′ ex ∆ex−1/bracketrightbigg ∆Teq (8) where we have relied on the fact that near the transition tem- perature, ∆ex≪kBTand made the approximation that δTe≪Tforlowlaserpower. Inthelastequation T0∆′ ex ∆ex≫1 nearTc,so onlythefirst termwill beconsidered. Thequantity ˙S(0),where˙S= 1/2(˙Nmaj−˙Nmin)/Ntotal, can be found by taking the derivative of ˙S(0)with respect to Te, since immediately after excitation the electron tempera- turehasincreased,butthespin temperature, T,hasnot. ˙S(0) =d˙S dTe/vextendsingle/vextendsingle/vextendsingle T=T0∆Teq=Nmaj N0τsf∆ex kBT∆Teq(9) Near the Curie temperature Nmaj∼Nmin∼1 2Ntotal. Usingthisapproximationsandequation(6), wefind: τM=/parenleftbigg∆′ ex ∆ex/parenrightbiggTcτsf 2(10) where∆′ existhederivativeof ∆exwithrespecttotemperature and∆ex∼(Tc−T)β, whereβis the critical exponent of the order parameter. Taking the derivative, we find ∆′ ex∼ −β(Tc−T)β−1,andthuscanwrite τM=βτsf 2/parenleftbiggTc Tc−T/parenrightbigg (11) Therefore τMis predicted to scale as 1/(Tc−T)near the transition temperature. A fit of T c∼148.8K is found for the datainFig 6. Note that detailed balance suggests that the demagnetiza- tion time scales as 1/tRnear the transition temperature re- gardlessoftheunderlyingmechanismofthedemagnetizatio n. Additionally,the critical exponentfoundis independento fβ. It should also be noted that the current situation, where the sample has been excited by a laser, is distinct from critical behavior as typically considered. In general, divergent ti me scales are linked to divergent length scales, but here excit a- tions of various length scales are not being excited. Instea d thelengthscale is alwayseffectivelyinfinite, havingbeen de- terminedby the laser spot size. τsfis plotted as a functionof temperature for the mean field value of β= 1/2, which has beenshowntobesuitableforSRO10,inFig. 8. τsfisrevealed to be approximately 200 fs and nearly constant as a function oftemperature.5 FIG.8. Spinfliptime at hightemperature Previous reports of conductivity in SRO give a scattering time of∼20 fs near the transition temperature11. A compar- ison of the spin flip time with the scattering time implies a probabilityof0.1thatascatteringeventsresultsinaspin flip. Thoughelectron-phononinteractionsare the most commonly considered source of demagnetization, as mentioned previ- ously, Eliot Yafet-like electron-electron coulomb scatte ring canalsoresultindemagnetization3. Thisisespeciallytruefor materials with strong spin orbit coupling, such as SRO. Ad- ditionally in SRO the interaction with the crystal field mean s that total spin is not conserved[Goodenough], so every elec - troninteractioncanperturbthespinstate. Having found a relationship between the demagnetization time and the spin flip time we would like to explore the rela- tionship between these parameters and the damping param- eter,α. Intuitively, the damping parameter should be pro- portional to the spin flip scattering rate, or inversely prop or- tionaltothe spinflip scatteringtime: α∼1/τsf. Elliot-Yafet type scattering dissipates energy from motion described by the LLG equation by disrupting the coherent, collective pre - cession of spins. Spins that have had their angular momen- tum changed through electron collisions must be pulled back into the precession through the exchange interaction, repr e-senting a transfer of energy away from the precessional mo- tion. These collision-mediatedspin-orbitcouplingeffec ts are thought to be the primary source of Gilbert-type damping in ferromagnets12. Again, this should be particularly true in a ferromagnetwith strongspinorbitcoupling. Combining the spin flip time and the damping parameter with Planck’s constant reveals an energy scale, E, given by theconditionthat: 1 α∼E /planckover2pi1τsf (12) Noting that the valuesfor αandτsffoundin figures 3 and 7, respectively, are approximately constant as a function o f temperature, this energy scale for SRO is ∼7 meV. The fun- damental energy scales applicable to the magnetic system in SRO aretheFermienergy,the exchangeenergy,andthecriti- caltemperature,thelasttwoofwhichareinterdependent. T he Fermi energy is orders of magnitude larger than 7 meV, but the energy associated with the critical temperature, kBTc∼ 13 meV, is of the same order. This suggests an underlying connectionbetween the critical temperature(and thus the e x- changeenergy),Gilbertdamping,andspinflip scattering. A relationshipsimilar to equation(12) hasbeenfoundpre- viously between τM(rather than τsf) andαby Koopmanset al. at lowtemperature: τM=1 4/planckover2pi1 kBTc1 α(13) Applying this equation to SRO at 5K yields τm∼30fs, which is unphysical since it is below the total scattering ra te of∼100fsatlowtemperature11. Whetherthefundamentalre- lationshipisbetweentransitiontemperatureandthedemag ne- tizationtimeorthespin-flipscatteringtimeremainsaques tion fora microscopicmodeltoresolve. ACKNOWLEDGMENTS This research is supported by the US Department of En- ergy, Office of Science under contract number DE-AC02- 05CH1123. 1E.Beaurepaire etal.,Phys. Rev. Lett. 76, 4250 (1996). 2B.Koopmans et al.,Nat.Mater. 9,259 (2009). 3M. Krauss et al.,Phys.Rev. B 80, 180407 (2009). 4T. Ogasawara et al.,Phys.Rev. Lett. 94, 087202 (2005). 5T. Kiseet al.,Phys.Rev. Lett. 85, 1986 (2000). 6M. Langner etal.,Phys. Rev. Lett. 102, 177601 (2009). 7G.Koster etal.,Appl. Phys.Lett. 73, 2920 (1998). 8B.Koopmans et al.,Phys.Rev. Lett. 95, 267207 (2005). 9N.Metropolis et al.,J.Chem. Phys. 21, 1087 (1953). 10D.Kim et al.,Phys.Rev. B 67, 100406 (2003).11J.S.Dodge et al.,Phys.Rev. Lett. 85, 4932 (2000). 12B. Heinrich, in Ultrathin Magnetic Structures III (Springer- Verlag,Berlin,Germany, 2005). 13J.Fabian, and S.Das Sarma, Phys.Rev. Lett. 81, 5624 (1998). 14P.Monod, andF.Beuneu, Phys.Rev. B 19, 911 (1979). 15X.Wanget al.,Phys.Rev. B. 74, 195118 (2006). 16D.L.Mills and S.M.Rezende, in Spin dynamics in confined mag- netic structures II , edited by B. Hillebrands and K. Ounadjela (Springer-Verlag,Berlin,Germany, 2003). 17M. Pickelet al.,Phys.Rev. Lett. 101, 066402 (2008). | 2010-08-03 | We report time-resolved Kerr effect measurements of magnetization dynamics in
ferromagnetic SrRuO3. We observe that the demagnetization time slows
substantially at temperatures within 15K of the Curie temperature, which is ~
150K. We analyze the data with a phenomenological model that relates the
demagnetization time to the spin flip time. In agreement with our observations
the model yields a demagnetization time that is inversely proportional to T-Tc.
We also make a direct comparison of the spin flip rate and the Gilbert damping
coefficient showing that their ratio very close to kBTc, indicating a common
origin for these phenomena. | Determination of the spin-flip time in ferromagnetic SrRuO3 from time-resolved Kerr measurements | 1008.0674v1 |
1 Enhancing Perpendicular Magnetic Anisotropy in Garnet Ferrimagnet by Interfacing with Few -Layer WTe 2 Guanzhong Wu1*, Dongying Wang1, Nishchhal Verma1, Rahul Rao2, Yang Cheng1, Side Guo1, Guixin Cao1, Kenji Watanabe3, Takashi Taniguchi4, Chun Ning Lau1, Fengyuan Yang1, Mohit Randeria1, Marc Bockrath1, and P. Chris Hammel1 1.Department of Physics, The Ohio State University, Columbus, OH, 43210, USA 2.Materials and Manufacturing Directorate, Air Force Research Laboratory, Wright -Patterson Air Force Base, Dayton, OH, 45433, USA 3. Research Center for Functional Materials, National Institute for Materials Science, 1 -1 Namiki, Tsukuba 305 -0044, Japan 4. International Center for Materials Nanoarchitectonics, Nationa l Institute for Materials Science, 1-1 Namiki, Tsukuba 305 -0044, Japan *wu.2314@osu.edu 2 Abstract: Engineering magnetic anisotropy in a ferro - or ferrimagnetic (FM) thin film is crucial in spintronic device. One way to modify the magnetic anisotropy is through the surface of the FM thin film. Here, we report the emergence of a perpendicular magnetic anisotropy (PMA) induced by interfacial interactions in a heterostructure comprised of a garnet ferrimagnet, Y 3Fe5O12 (YIG), and the low -symmetry, high spin orbit coupling (SOC) transition metal dichalcogenide, WTe 2. At the same time, we also observed an enhancement in Gilbert damping in the WTe 2 covered YIG area. Both the magnitude of interface -induced PMA and the Gilb ert damping enhancement have no observable WTe 2 thickness dependence down to single quadruple -layer, indicating that the interfacial interaction plays a critical role. The ability of WTe 2 to enhance the PMA in FM thin film, combined with its previously rep orted capability to generate out -of-plane damping like spin torque, makes it desirable for magnetic memory applications. Key words: perpendicular magnetic anisotropy, magnetic resonance force microscope, transition metal dichalcogenides, ferrimagnetic i nsulator 3 Perpendicular magnetic anisotropy (PMA) in a ferromagnetic thin film is of great interest in spintronics research and application s. Ferromagnetic nano -element s with PMA overcome their shape anisotropy , greatly ease the memory cell size reduction and improves memory retention . These exceptional properties, improving the performance of magnetic devices , make PMA highly desirable for magnetic memory application s. PMA becomes even more important in the recent development of solid state magnetic random -access memory (MRAM) since it allows MRAM to have lower switching current and faster switching speed compare d to in-plane magnetized materials 1, 2. Magnetic storage devices generally rely on metallic magnetic material s due to their robust electrical response . Interfacial magnetic anisotropy plays a critical role in generat ing PMA in metallic ferromagnet s. When interfacing with a nonmagnetic material (NM), electron orbital angular momentum of the magnetic ions at the ferromagnet surface will be modified, in some cases enabling strong covalent bonding, resulting in distinct magnetic properties compare d to the single layer 3-6. However, spintronics devices made of metallic magnetic materials are inherently energy consumptive due to resistive losses. Recently, complex oxide ferro - or ferrimagnet insulator s (FMI) have attracted substantial interest due to their ability to transport spin excitation s with low dissipation 7. Inducing PMA in FMIs naturally becomes an important topic both for scientific and technologic al reasons . Several successful route s to achiev ing PMA in FMIs ha ve been reported using bulk intrinsic anisotropy 8 or lattice strain 9-12. But in most experiments, the sign of the resulting interfacial anisotropy in FMI/NM heterostructures is such as to enhance the easy-plane anisotropy 13-15. Only one recent experiment has shown the possibility of generating interfacial PMA, and this was attributed to topological surface states 16. Nevertheless, t hese results demonstrate the possibility of controlling magnetic anisotropy through interfac ial interaction s in 4 FMI/NM heterostructures . Here, we report a study on YIG/WTe 2/hBN heterostructures, which shows that when interfacing with a low symmetry nonmagnetic van der Waals material , WTe 2, an additional interfac e-induced PMA (iPMA) term emerges in the magnetic anisotropy of the YIG thin film . The absence of topological surface states at room temperature in WTe 2 17, 18 forces us to seek an explanation for our observation of enhanced PMA that is distinct from that proposed for top ological insulator/YIG bilayers 16. We therefore turn to an analysis of the broken symmetries in WTe 2. We point out that low symmetry WTe 2 has recently shown the capability of generating both in -plane and out -of-plane spin polarization in charge -spin conversion experiments 19-22. It also enables field-free switching of PMA magnet ic material , which ease s the application of PMA material s in MRAM application s 23-25. Ferrimagnetic insulator YIG is of significant research interest in spintronics due to its exceptionally low Gilbert damping 26, which describes the relaxation rate of magnetization precession . And 1T’-WTe 2 is a semi -metallic transition metal dichalcogenide (TMD) layered material with strong SOC 27, 28. The crystal structure of 1T’-WTe 2 lacks twofold rotational symmetry about the c -axis (Fig. 1a). The only symmetry in the WTe 2 crystal lattice ab plane is the mirror symmetry about the bc plane 29. This u nique symmetry breaking allows out-of-plane damping -like torque to be generated 30, 31, enabling efficient switching of the out-of-plane magnetization of the adjacent magnetic material 24. A 20nm thick YIG thin film used in our experiment is epitaxially grown on (111) -oriented Gd3Ga5O12 (GGG) substrate by off -axis sputtering 32. WTe 2 flakes are then mechanically exfoliated from a flux-grown crystal, and dry transferred on to the clean top YIG surface without touching any other substances. This whole process is carried out in an Ar -filled glove box with <0.1 ppm of H 2O and 5 O2 to protect the flakes from degradation and ensure the clean liness of the YIG/WTe 2 interface. We employ hexagonal boron nitride ( hBN ) encapsulation to protect the WTe 2 flakes from oxidation after being removed from the glove box. We make two samples and focus on the data taken from sample 1 in the main text. The raw data taken from sample 2 can be found in Supporting Information Fig. S 2. Fig. 1 Crystal structure of WTe 2 and sample schematic . a) Crystal lattice structure of WTe 2 viewed from the top along the c-axis and looking from the side along the a-axis. The black dashed box in the side view indicates a monolayer of WTe 2. b) Schematic of the ferromagnetic resonance force microscope. RF excitation is generated by a stripline underneath the sample , where the hBN encapsulation is not shown . The region of localized mode is shown as a yellow dot adjacent to the WTe 2 flake, and the probe magnetic moment is shown as a yellow arrow on the particle. The cantilever oscillation is detected by a fiber laser interferometer. Figure 2a shows an optical image of the sample 1. Due to the small lateral size of the exfoliate d WTe 2 and hBN flake s having length scale s of 10 μm, we use a home -built ferromagnetic resonance force microscope (FMRFM) to measure the local ferromagnetic resonance (FMR) signal. FMRFM is a sensitive technique to detect the local magnetic properties with high spatial and spectral resolution 33. In our FMRFM, the external magnetic field 𝐻⃗⃗ ext is aligned perpendicular to the sample plane. The cantilever tip holds a high coercivity SmCo 5 magnetic particle , whose moment is magnetized in the 6 direction opposite to 𝐻⃗⃗ ext to create a magnetic field well . The field well supports a set of localized standing spin wave modes (LMs). During the measurement, we excite spin precession uniformly by a stripline underneath the sample at a fixed RF frequency (2 GHz) and sweep the magnetic field. The resonance of each LM generate s a stray field, whic h can then be detected by the SmCo 5 magnetic particle attached on the cantilever through their magnetic dipole -dipole interaction (Fig. 1b). During the measurement, we keep the probe -to-sample separation around 4 μm. The operation of FMRFM is described in detail in Ref s. 34-36. For reference, w e separate a region of YIG that does not contain WTe 2/hBN heterostructures and measur e its Gilbert damping using broadband FMR. To eliminate two - magnon scattering, w e perform broadband FMR in the out -of-plane field geometry. The FMR linewidth as a function of frequency measured on bare YIG (sample 1) shows a linear dependence (Fig. 2b), from which we can extract the Gilbert damping of bare YIG 𝛼YIG=1.05×10−3. We also confirm that the WTe 2 used in the experiment is indeed the 1T’ phase through polarized Raman measurements. The polarization angle dependence of the Raman peak at 212 cm-1 (spectrum is shown in Fig. S4) exhibits minimum intensity when the excitation laser polarization is along the crystallographic a axis of WTe 2 37 as shown by the polar plot in Fig. 2c and Raman intensity plot in Fig. 2d . We find t he position of the YIG/WTe 2/hBN heterostructure with the assist ance of magnetic alignment markers (Fig. 2a) . Figure 2e shows t wo raw FMRFM scans taken in the region of YIG/hBN and YIG/WTe 2/hBN , indicated by the blue and the red dot in Fig. 2a, respectively , which reveals the change in FMRFM spectra at two different location s. Here we focus on the 𝑛=1 LM because it has the mode radius of around 1 μm and gives the highest spatial resolution. Higher order modes have increasing mode radius and therefore, detect less local ized magnetic properties. This is the reason why 7 the quasi -uniform mode at ~ 3325 Oe does not show obvious change in resonance field or signal amplitude. We further take a line scan across the edge of WTe 2 flake (Fig. 2 f) to resolve the spatial evolution of FMRFM spectra . The line scan in Fig. 2f (along the dashed line shown in Fig. 2a) shows three main features : first, the magnitude of the LM resonance signal is reduced in the YIG/WTe 2/hBN region compare d to the YIG/hBN region; second, the LM resonance field for all LMs is decreased by ~40 Oe in the YIG/WTe 2/hBN region; third , the LMs show complex splitting and crossing when the probe is close to the boundary (−5 μm<𝑋<10 μm). Fig. 2 FMRFM and Raman measurement data . a) An optical micrograph of the YIG/WTe 2/hBN heterostructure under study . WTe 2 crystal a and b axis are labeled. b) Broadband FMR measurement of the frequency -dependent linewidth of the YIG thin film. The measurement is done on the same piece of YIG used to make sample shown in Fig. 1b. c) Polar plot of the 21 2 cm-1 peak Raman intensity. Angle denote s the relative angle between the measurement laser polarization and the WTe 2 a axis. d) 2D intensity plot showing Raman peak intensities versus polarization angle . e) FMRFM spectra, one over the YIG/hBN region (blue line) and the second over the YIG/WTe 2/hBN region (red line); these locations are indicated by the blue and red dot s in Fig. 2a respectively . f) Color plot of field -dependence 8 FMRFM scans as a function of position along the trace indicated by the black dashed line in Fig. 2a. A constant background is subtracted to show only the signal from the several LM resonance s. In the following, we will explain the origin of the three observed effects using spin pumping and magnetic anisotropy. The first effect , i.e. signal reduction in the YIG/WTe 2/hBN area relative to the YIG/hBN area, is the result of enhanced relaxation due to spin pumping from YIG to WTe 2 38. The 𝑛= 1 LM resonance signal amplitude ∆𝐴 is inversely proportional to the square of Gilbert damping , 𝛼2. We determine the Gilbert damping constant 𝛼 for YIG/WTe 2/hBN using 𝛼YIG/WTe2/hBN= 𝛼YIG/hBN×√∆𝐴YIG/hBN∆𝐴YIG/WTe2/hBN ⁄ (see Ref. 39), where 𝛼YIG/hBN is assumed to be the same as 𝛼YIG=1.05×10−3 due to the low SOC and insulating character of hBN . The second effect is the decrease of 𝑛=1 LM resonance field 𝐻r,1 by ~40 Oe . And the third effect is splitting and crossing of complex modes in the region −5 μm<𝑋<10 μm. The second and the third effects are due to an abrupt change of uniaxial anisotropy across the boundary separating the YIG/WTe 2/hBN and YIG/hBN regions 15. Here, the uniaxial anisotropy refers to the magnetic free energy depends on the angle between magnetization and sample normal ℱu=−𝐾u𝒎z2, where 𝒎z is the component of magnetization unit vector in the direction normal to sample plane and 𝐾u is the uniaxial anisotropy constant specific to sample and depends on the total interaction in the sample . When 𝐾u is positive, ℱu is called to be of PMA type, on the other hand, if 𝐾u is negative, ℱu is called to be of easy -plane type. This uniaxial anisotropy will lead to an effective uniaxial magnetic field 𝑯u=−𝜕ℱu𝜕𝑴⁄ , where 𝑴 is the magnetization . And therefore, a change in 𝐾u can modify the resonance field in a FMR measurement . In FMRFM spatial mapping, a n abrupt change in 𝐾u spatially could disturb the LM and lead to mode splitting and crossing as described in Ref. 15. Moreover, i n striking contrast to the previously studied 9 YIG/Au interface 15, which result s in a 32 Oe increase of 𝐻r,1 due to the enhanced easy -plane anisotropy, the observed decrease of 𝐻r,1 indicates that the WTe 2 overlayer induces an iPMA in YIG. We note that the magnitude of the shift in 𝐻r,1 is comparable to the easy -plane anisotropy induced by a heavy metal 15, 40 or the iPMA generated by topological surface state 16 on garnet ferrimagnetic material . In order to probe the global effect of a WTe 2 overlayer on YIG, we spatial ly map 𝐻r,1 using the 𝑛=1 LM. Figure 3a present s an optical image of WTe 2 flakes on a Si/SiO 2 (285nm ) substrate , where different c olors of WTe 2 flakes indicat e different WTe 2 thickness es. Figure 3b and 3 c show spatial maps of magnetic properties in the region enclosed by the black dashed rectangle in Fig. 3a . We acquire the maps using the procedure described in Ref. 39, i.e., simultaneously measuring spatial variation of the magnetic anisotropy and Gilbert damping using the 𝑛=1 LM resonance field 𝐻r,1 and signal amplitude ∆𝐴. The entire WTe 2-covered area show s uniformly lower ed 𝐻r,1 and increased Gilbert damping relative to the area without WTe 2. In Fig. 3c, despite the not great signal to noise ratio in damping imaging, the re is a clear Gilbert damping enhancement in WTe 2-covered area . The averaged Gilbert damping of YIG in WTe 2-covered area is 𝛼̅YIG/WTe2/hBN≈1.30×10−3, about 24% higher than 𝛼YIG. We note that due to the slight relative tilting of the scan plane and the sample plane, there is a color shift in Fig. 3b that might conceal the contrast difference in different WTe 2 thickness region . Therefore, to study the WTe 2 thickness dependence, we will show fine line scans across edge s of flakes having different WTe 2 thickness es. 10 Fig. 3 Two -dimensional FMRFM scan resolving the spatial variation of magnetic anisotropy and Gilbert damping. a ) Optical micrograph showing the color contrast of different thickness WTe 2 flake s (ranging from 4.7 nm to 44.8 nm) on Si/SiO 2(300 nm). Black dashed box outlines the FMRFM scanned area for 2D mapping. b) 2D map of the 𝑛=1 LM resonance field . The d ashed line s labeled 1 -4 correspond to the four line-scans shown in Fig. S1a-S1d. c) 2D mapping of the Gilbert damping extracted from the 𝑛=1 LM resonance peak amplitude. 11 Next, we want to understand what gives rise to the PMA in WTe 2/YIG. We rule out the effect induced by a modification of the gyromagnetic ratio by showing the resonance field shift across the WTe 2 edge does not depend on RF excitation frequency (See Fig. S3). We also exclude a strain induced effect given the absence of an epitaxial relation and the weakness of the van der Waals interaction between YIG and WTe 2. We further note that we can ignore the role of topological surface states 16 in our analysis; they are not relevant for our room temperature experiment since WTe 2 is a topological Weyl semimetal only below 100 K 17, 18. We show how an analysis based on symmetry and the nature of the interfacial SOC , generalizing the theory in Ref. 41, gives insight into the PMA observed in our experiment. This will also help us understand why the easy-axis anisotropy we observe in WTe 2/YIG is so different from the results of Ref. 13, 15 on YIG interfaces with a dozen different metallic and semiconducting materials, all of which exhibit interface -induced easy-plane anisotropy, as is predicted by theory 41. YIG is a ferrimagnetic Mott insulator, with two inequivalent Fe -sites coupled via antiferromagnetic (AFM) superexchange interactions. We focus on how interfacial SOC impacts AFM superexchange in YIG and show that it leads to a very specific form of the mag netic anisotropy that is governed by the direction of the effective B-field (see Supporting Information for a details). Before turning to WTe 2/YIG, it is useful to first consider the simpler case when the only broken symmetry is the mirror plane defined b y the interface. The abrupt change in lattice potential then results in an effective electric field that points normal to the interface, which in turn leads to an effective magnetic field in the rest frame of the electron that couples to its spin. Since t he E-field points normal to 12 the interfacial plane in which the electron moves, the resulting B-field arising from SOC lies within the interfacial plane. As we show in the SI, this leads to a SOC -induced correction to AFM superexchange that necessarily lead s to an easy-plane anisotropy. In the case of WTe 2/YIG, however, when there are additional broken symmetries. Not only does the interface break inversion symmetry , but the crystal structure of WTe 2 itself breaks in -plane inversion symmetry . The electric field is now no longer normal to the interface, and the effective B-field arising from SOC necessarily has an out -of-plane component, as shown in Fig S5b in SI. Thus, we see why the lower symmetry of WTe 2/YIG can naturally result in an easy-axis or perpendicular magnetic anisotropy (PMA); see Supporting Information for details. We note that the lack of two -fold rotational symmetry in the ab plane in WTe 2 that plays a critical role in our understanding of PMA in WTe 2/YIG, has also been pointed out be crucial for the out - of-plane damping -like torque in WTe 2/Permalloy30. We note, however, that the out -of-plane damping - like torque necessarily involves current flow in WTe 2, while the PMA is an equilibrium property of the system independent of current flow. We further demonstrate the interfacial origin of the observed effect by studying the influence of WTe 2 thickness. We show four line-scans , labeled in Fig. 3b, across the edges of WTe 2 with different thickness es, ranging from 4. 7 nm to 44.8 nm . From these four line -scans, we extract the 𝑛=1 LM resonance field 𝐻r,1 and the 𝑛=1 LM resonance signal amplitude ∆𝐴. Figures S1a-d in the Supporting Information show the evolution of 𝐻r,1 and ∆𝐴 along the traces labeled correspondingly . The thickness of WTe 2 at each measurement location is later measured using atomic force microscop y. From these 13 line-scans , we choose the region s where the probe is far away from the edge of WTe 2 so that the magnetic propert ies are uniform, to obtain spatial average s of 𝐻r,1 and ∆𝐴, which are denoted 𝐻̅r,1,YIG/hBN and ∆𝐴̅̅̅̅YIG/hBN in the YIG/hBN region , and 𝐻̅r,1,YIG/WTe2/hBN and ∆𝐴̅̅̅̅YIG/WTe2/hBN in the YIG/WTe 2/hBN region , respectively . We further extract the 𝑛=1 LM resonance field difference between two regions using ∆𝐻r,1=𝐻̅r,1,YIG/hBN−𝐻̅r,1,YIG/WTe2/hBN, as well as the Gilbert damping difference using ∆𝛼=𝛼YIG×(√∆𝐴̅̅̅̅YIG/hBN ∆𝐴̅̅̅̅YIG/WTe2/hBN ⁄ −1) as a function of the WTe 2 thickness . We note that the hBN overlayer does not change the Gilbert damping in YIG . The summarized results containing the data from both sample 1 and sample 2 are shown in Fig s. 4a and 4 b. Raw data from sample 2 can be found in Supporting Information Fig. S 2. The thinnest WTe 2 acquired in the experiment is 3.2nm from sample 2, which is approximately the thickness of a quadruple -layer WTe 2. Figures 4a and 4 b indicate that both ∆𝐻r,1 and ∆𝛼 have almost no WTe 2 thickness dependence . There is a small sample -to-sample variation possibly due to different YIG/WTe 2 interfacial quality . The change of 𝑛=1 LM resonance field, ∆𝐻r,1, is as large as ~38 Oe even when the WTe 2 thickness approaches the quadruple -layer thickness . This indicates that the modification of magnetic anisotropy is due to the YIG/WTe 2 interfac ial interaction , with no bulk contribution. For the increase of Gilbert damping ∆𝛼, no obvious thickness dependence is observed when comparing the data from the same sample. In sample 2, the Gilbert damping enhancement due to the quadruple -layer WTe 2 has almost the same value as the 50 nm thick WTe 2 flake, indicating that no thickness dependence of spin pumping can be resolved from our measurement. There are two possible interpretations of these results . First, if the 14 spin current injected into WTe 2 is mainly relaxed due to spin relaxation in the bulk, then the experimental result is a demonstration of ultra -short spin diffusion length along the c axis38, smaller or comparable to the thinnest WTe 2 flake (3.2 nm), employed in this experiment . It is much smaller than the 8nm spin diffusio n length in the in-plane direction measured using inverse spin Hall effect 22. Note that due to the chang e in mo bility and the metal -insulator transition in few layer WTe 2 when its thickness reduces 42, the spin diffusion length approximated here could be inaccurate . Alternatively , it is possible that the spin relaxation is primarily due to the interfacial SOC induced by inversion symmetry breaking at the interface and in the WTe 2 crystal lattice. In this case, the Gilbert damping enhancement will have no WTe 2 thickness dependence. 15 Fig. 4 WTe 2 thickness dependence of resonance field and damping enhancement . a) 𝐻r,1 in the YIG/hBN and YIG/WTe 2/hBN regions are averaged respectively to get 𝐻̅r,1,YIG and 𝐻̅r,1,YIG/WTe2, and ∆𝐻r,1=𝐻̅r,1,YIG−𝐻̅r,1,YIG/WTe2. b) ∆𝛼 as a function of WTe 2 thickness , and ∆𝛼=𝛼YIG/WTe2−𝛼YIG where 𝛼YIG is the Gilbert damping of bare YIG measured using broadband FMR for each sample , and 𝛼YIG/WTe2=𝛼YIG×(√∆𝐴̅̅̅̅YIG/hBN ∆𝐴̅̅̅̅YIG/WTe2/hBN ⁄ ). In conclusion, we have shown that the YIG/WTe 2 interface plays a critical role in both interfacial magnetic anisotropy and spin relaxation , making WTe 2 a promising material in magnetic memory 16 application s. Combining the iPMA created by WTe 2 with the out-of-plane spin orbit torque generated by flowing a charge current along the a axis of WTe 2, one can possibly achieve field -free switching of a PMA magnetic cell for magnetic memory application s. It will improv e the scalability , reduc e the power consumption and increas e operation speed of magnetic solid -state devices . Our result reveals new possibilities in selecting materials and designing spintronic devices. For example, one can consider other materials with low lattice symmetry and strong SOC to induce larger PMA type interfacial ani sotropy in FMIs. To achieve a fully PMA material, one could utilize thinner FMIs to magnify the role of iPMA. Moreover, interfacial SOC also plays an important role in generat ing topologically protected magnetic textures in the FMIs 43. These findings will motivate further research to reveal the fundamental physics arising at the interface between FMIs and nonmagnetic materials. Data availability: The data generated by the present study are available from the corresponding author on request. Supporting Information: A description of raw data on WTe 2 thickness dependence, a FMRFM measurement on a second sample, a FMRFM measurement at different RF frequency, a description of polarized Raman measurement result, and a detailed illustration of impact of broken mirror reflection symmetries on the magnetic anisotropy. Ackno wledgements: 17 This work was primarily supported by the Center for Emergent Materials: an NSF MRSEC under award number DMR -2011876 (GW, NV , YC, SG, FY , MR and PCH) . KW and TT acknowledge support from the Elemental Strategy Initiative conducted by the MEXT, Japan (Grant Number JPMXP0112101001) and JSPS KAKENHI (Grant Numbers 19H05790, 20H00354 and 21H05233). DW, GC, CNL, and MB are supported by NSF under award DMR -2004801. We gr atefully acknowledge N. Trivedi for insightful discussions. Fabrication and some characterization were performed in the Ohio State University NanoSystems Laboratory. 18 Methods: Sample fabrication Our YIG/WTe 2/hBN heterostructure was prepared by means of dry transfer and stacking 44. hBN crystals were mechanically exfoliated under ambient conditions onto SiO 2/Si substrates (285 nm thick SiO 2). 20- 40 nm thick hBN flakes were identified under an optical microscope and used for the capping lay er for the stack. The hBN was picked up using a polymer -based dry transfer technique and then moved into an Ar-filled glove box with oxygen and water level below 0.1 ppm. Flux -grown WTe 2 crystals 45 were exfoliated inside the glove box and flakes with different thicknesses were optically identified and quickly picked up with the capping hBN layer then transferred to the YIG substrate. Finally, we removed the fully encapsulated sample from the glove b ox and performed the e -beam lithography and metallization (Ni/Au) step for alignment in our ferromagnetic resonance force microscope (FMRFM). Polarized Raman measurement Polarized Raman spectra from the WTe 2 sample were collected using 633 nm excitation w avelength in an inVia Renishaw Raman microscope. The sample was loaded onto the microscope stage and positioned in such a way that the long edge of the flake was aligned parallel to the laser polarization ( θ = 0°). In this configuration, the incident illu mination is polarized vertically coming out of the laser and is aligned with the long axis of the WTe 2 flake. The polarization of the incident laser was rotated from 0 to 360° by 10° increments using a polarization rotator, while an analyzer was set to onl y allow vertically polarized light to enter the spectrometer. Raman spectra were collected at each polarization for 3 acquisitions with a 20 s time per acquisition. The laser power was set to 0.5 mW at the sample to avoid any damage by heating. Followin g spectral collection, the (baseline corrected) integrated intensities under each peak were calculated to make the contour plots and polar plots in Fig. 2c and 2d. FMRFM measurement and signal fitting Our FMRFM perform s local ly measures FMR at room temper ature in vacuum. The cantilever has natural frequency of ~18 KHz, spring constant of 0.2 N/m and Q factor of ~20000, resulting in force detection sensitivity of 10-15 N/Hz1/2. The SmCo 5 magnetic particle attached on the cantilever has a magnetic moment of ~4 nemu. When a LM is on resonance, the local reduction of magnetization in out - of-plane direction will generate a stray field, which will couple the altered magnetization to the magnet ic tip thus changing the cantilever oscillation amplitude and frequency. The change in cantilever oscillation is detected optically by laser fiber interferometry. Different LMs have different mode radi i. For this microscopy study, we focus on 𝑛=1 LM since it gives the highest spatial resolution. The 𝑛=1 LM resonance peak is fit to a Lorentzian line shape, from which the peak position and peak height are extracted, which correspond to the resonance field 𝐻r,1 and signal amplitude ∆𝐴. 19 References: (1) Dieny, B.; Chshiev, M. Perpendicular magnetic anisotropy at transition metal/oxide interfaces and applications. Rev. Mod. Phys. 2017, 89, (2), 025008. (2) Lee, K. S.; Lee, S. W.; Min, B. C.; Lee, K. J. Threshold current for switching of a perpendicula r magnetic layer induced by spin Hall effect. Appl. Phys. Lett. 2013, 102, (11), 112410. (3) Carcia, P. F.; Meinhaldt, A. D.; Suna, A. Perpendicular Magnetic -Anisotropy in Pd/Co Thin -Film Layered Structures. Appl. Phys. Lett. 1985, 47, (2), 178 -180. (4) Ikeda, S.; Miura, K.; Yamamoto, H.; Mizunuma, K.; Gan, H. D.; Endo, M.; Kanai, S.; Hayakawa, J.; Matsukura, F.; Ohno, H. A perpendicular -anisotropy CoFeB -MgO magnetic tunnel junction. Nat. Mater. 2010, 9, (9), 721 -724. (5) Yang, H. X.; Chen, G.; Cotta, A. A. C.; N'Diaye, A. T.; Nikolaev, S. A.; Soares, E. A.; Macedo, W. A. A.; Liu, K.; Schmid, A. K.; Fert, A.; Chshiev, M. Significant Dzyaloshinskii -Moriya interaction at graphene -ferromagnet interfaces due to the Rashba effect. Nat. Mater. 2018, 17, (7), 60 5-609. (6) Zeper, W. B.; Greidanus, F. J. A. M.; Carcia, P. F. Evaporated Co/Pt Layered Structures for Magneto - Optical Recording. Ieee Transactions on Magnetics 1989, 25, (5), 3764 -3766. (7) Brataas, A.; van Wees, B.; Klein, O.; de Loubens, G.; Viret, M. Spin insulatronics. Phys Rep 2020, 885, 1 -27. (8) Li, P.; Liu, T.; Chang, H. C.; Kalitsov, A.; Zhang, W.; Csaba, G.; Li, W.; Richardson, D.; DeMann, A.; Rimal, G.; Dey, H.; Jiang, J.; Porod, W.; Field, S. B.; Tang, J. K.; Marconi, M. C.; Hoffmann, A.; Mr yasov, O.; Wu, M. Z. Spin -orbit torque -assisted switching in magnetic insulator thin films with perpendicular magnetic anisotropy. Nat. Commun. 2016, 7, 1-8. (9) Tang, C.; Sellappan, P.; Liu, Y . W.; Xu, Y . D.; Garay, J. E.; Shi, J. Anomalous Hall hysteres is in Tm3Fe5O12/Pt with strain -induced perpendicular magnetic anisotropy. Phys. Rev. B 2016, 94, (14), 140403. (10) Soumah, L.; Beaulieu, N.; Qassym, L.; Carretero, C.; Jacquet, E.; Lebourgeois, R.; Ben Youssef, J.; Bortolotti, P.; Cros, V .; Anane, A. Ult ra-low damping insulating magnetic thin films get perpendicular. Nat. Commun. 2018, 9, 1-6. (11) Li, G.; Bei, H.; Su, J.; Zhu, Z. Z.; Zhang, Y .; Cai, J. W. Tunable perpendicular magnetic anisotropy in epitaxial Y3Fe5O12 films. Apl Materials 2019, 7, (4), 041104. (12) Fu, J. B.; Hua, M. X.; Wen, X.; Xue, M. Z.; Ding, S. L.; Wang, M.; Yu, P.; Liu, S. Q.; Han, J. Z.; Wang, C. S.; Du, H. L.; Yang, Y . C.; Yang, J. B. Epitaxial growth of Y3Fe5O12 thin films with perpendicular magnetic anisotropy. Appl. Phys. Le tt. 2017, 110, (20), 202403. (13) Lee, A. J.; Ahmed, A. S.; McCullian, B. A.; Guo, S. D.; Zhu, M. L.; Yu, S. S.; Woodward, P. M.; Hwang, J.; Hammel, P. C.; Yang, F. Y . Interfacial Rashba -Effect -Induced Anisotropy in Nonmagnetic - Material -Ferrimagnetic -Insulator Bilayers. Phys. Rev. Lett. 2020, 124, (25), 257202. (14) Tang, C.; Song, Q.; Chang, C. Z.; Xu, Y . D.; Ohnuma, Y .; Matsuo, M.; Liu, Y . W.; Yuan, W.; Yao, Y . Y .; Moodera, J. S.; Maekawa, S.; Han, W.; Shi, J. Dirac surface state -modulated spin dynamics in a ferrimagnetic insulator at room temperature. Science Advances 2018, 4, (6), eaas8660. (15) Wu, G. Z.; White, S. P.; Ruane, W. T.; Brangham, J. T.; Pelekhov, D. V .; Yang, F. Y .; Hammel, P. C. Local measurement of interfacial interactions using ferrom agnetic resonance force microscopy. Phys. Rev. B 2020, 101, (18), 184409. 20 (16) Liu, T.; Kally, J.; Pillsbury, T.; Liu, C. P.; Chang, H. C.; Ding, J. J.; Cheng, Y .; Hilse, M.; Engel - Herbert, R.; Richardella, A.; Samarth, N.; Wu, M. Z. Changes of Magnetism in a Magnetic Insulator due to Proximity to a Topological Insulator. Phys. Rev. Lett. 2020, 125, (1), 017204. (17) Li, P.; Wen, Y .; He, X.; Zhang, Q.; Xia, C.; Yu, Z. M.; Yang, S. Y . A.; Zhu, Z. Y .; Alshareef, H. N.; Zhang, X. X. Evidence for topological type-II Weyl semimetal WTe2. Nat. Commun. 2017, 8, 1-8. (18) Lv, Y . Y .; Li, X.; Zhang, B. B.; Deng, W. Y .; Yao, S. H.; Chen, Y . B.; Zhou, J.; Zhang, S. T.; Lu, M. H.; Zhang, L.; Tian, M. L.; Sheng, L.; Chen, Y . F. Experimental Observation of Anisotropic Adler -Bell- Jackiw Anomaly in Type -II Weyl Semimetal WTe1.9 8 Crystals at the Quasiclassical Regime. Phys. Rev. Lett. 2017, 118, (9), 096603. (19) Shi, S. Y .; Li, J.; Hsu, C. H.; Lee, K.; Wang, Y .; Li, Y .; Wang, J. Y .; Wang, Q. S.; Wu, H.; Zhang, W. F.; Eda, G.; Liang, G. C. A.; Chang, H. X.; Yang, Y . O. O. Observ ation of the Out -of-Plane Polarized Spin Current from CVD Grown WTe2. Adv Quantum Technol 2021, 4, (8). (20) Shi, S. Y .; Liang, S. H.; Zhu, Z. F.; Cai, K. M.; Pollard, S. D.; Wang, Y .; Wang, J. Y .; Wang, Q. S.; He, P.; Yu, J. W.; Eda, G.; Liang, G. C.; Yang, H. All -electric magnetization switching and Dzyaloshinskii - Moriya interaction in WTe2/ferromagnet het erostructures. Nature Nanotechnology 2019, 14, (10), 945 - 949. (21) Zhao, B.; Karpiak, B.; Khokhriakov, D.; Johansson, A.; Hoque, A. M.; Xu, X. G.; Jiang, Y .; Mertig, I.; Dash, S. P. Unconventional Charge -Spin Conversion in Weyl -Semimetal WTe2. Adv Mater 2020, 32, (38). (22) Zhao, B.; Khokhriakov, D.; Zhang, Y .; Fu, H. X.; Karpiak, B.; Xu, X. G.; Jiang, Y .; Yan, B. H.; Dash, S. P.; Anamul. Observation of charge to spin conversion in Weyl semimetal WTe2 at room temperature. Phys Rev Res 2020, 2, (1). (23) Fukami, S.; Anekawa, T.; Zhang, C.; Ohno, H. A spin -orbit torque switching scheme with collinear magnetic easy axis and current configuration. Nature Nanotechnology 2016, 11, (7), 621 -625. (24) Kao, I. -H.; Muzzio, R.; Zhang, H.; Zhu, M.; Gobbo, J.; Weber, D.; Rao, R.; Li, J.; Edgar, J. H.; Goldberger, J. E.; Yan, J.; Mandrus, D. G.; Hwang, J.; Cheng, R.; Katoch, J.; Singh, S. Field -free deterministic switching of a perpendicularly polarized magnet using unconventional spin -orbit torques in WTe2. arXiv 2020 . https://arxiv.org/ftp/arxiv/papers/2012/2012.12388.pdf (accessed Jan 22, 2022) (25) Shin, I.; Cho, W. J.; An, E. -S.; Park, S.; Jeong, H. -W.; Jang, S.; Baek, W. J.; Park, S. Y .; Yang, D. -H.; Seo, J. H.; Kim, G. -Y .; Ali, M. N.; Choi, S. -Y .; Lee, H. -W.; Kim , J. S.; Kim, S.; Lee, G. -H. Spin -orbit torque Switching in an All -Van der Waals Heterostructure. arxiv 2021 . https://arxiv.org/ftp/arxiv/papers/2102/2102.09300.pdf (accessed Jan 22, 2022) (26) Gallagher, J. C.; Yang, A. S.; Brangham, J. T.; Esser, B. D.; White, S. P.; Page, M. R.; Meng, K. Y .; Yu, S. S.; Adur, R.; Ruane, W.; Dunsiger, S. R.; McComb, D. W.; Yang, F. Y.; Hammel, P. C. Exceptionally high magnetization of stoichiometric Y3Fe5O12 epitaxial films grown on Gd3Ga5O12. Appl. Phys. Lett. 2016, 109, (7), 072401. (27) Ali, M. N.; Xiong, J.; Flynn, S.; Tao, J.; Gibson, Q. D.; Schoop, L. M.; Liang, T.; Haldolaarachchige, N.; Hirschberger, M.; Ong, N. P.; Cava, R. J. Large, non -saturating magnetoresistance in WTe2. Nature 2014, 514, (7521), 205 -208. (28) Tang, S. J.; Zhang, C. F.; Wong, D.; Pedramrazi, Z.; Tsai, H. Z.; Jia, C. J.; Moritz, B.; Claassen, M.; Ryu, H.; Kahn, S.; Jiang, J.; Yan, H.; Hashimoto, M.; Lu, D. H.; Moore, R. G.; Hwang, C. C.; Hwang, C.; 21 Hussain, Z.; Chen, Y . L.; Ugeda, M. M.; Liu, Z .; Xie, X. M.; Devereaux, T. P.; Crommie, M. F.; Mo, S. K.; Shen, Z. X. Quantum spin Hall state in monolayer 1T ' -WTe2. Nat. Phys. 2017, 13, (7), 683 -+. (29) Kang, K. F.; Li, T. X.; Sohn, E.; Shan, J.; Mak, K. F. Nonlinear anomalous Hall effect in few -layer WTe2. Nat. Mater. 2019, 18, (4), 324 -+. (30) MacNeill, D.; Stiehl, G.; Guimaraes, M.; Buhrman, R.; Park, J.; Ralph, D. Control of spin –orbit torques through crystal symmetry in WTe 2/ferromagnet bilayers. Nat. Phys. 2017, 13, (3), 300. (31) MacNeill, D.; Stiehl, G. M.; Guimaraes, M. H. D.; Reynolds, N. D.; Buhrman, R. A.; Ralph, D. C. Thickness dependence of spin -orbit torques generated by WTe2. Phys. Rev. B 2017, 96, (5), 054450. (32) Yang, F. Y .; Hammel, P. C. FMR -driven spin pumping in Y3Fe5O12 -based structures. Journal of Physics D -Applied Physics 2018, 51, (25), 253001. (33) Lee, I.; Obukhov, Y .; Xiang, G.; Hauser, A.; Yang, F.; Banerjee, P.; Pelekhov, D. V .; Hammel, P. C. Nanoscale scanning probe ferromagnetic resonance imaging us ing localized modes. Nature 2010, 466, (7308), 845. (34) Adur, R.; Du, C.; Manuilov, S. A.; Wang, H.; Yang, F.; Pelekhov, D. V .; Hammel, P. C. The magnetic particle in a box: Analytic and micromagnetic analysis of probe -localized spin wave modes. J. Appl. Phys. 2015, 117, (17), 17E108. (35) Du, C.; Lee, I.; Adur, R.; Obukhov, Y .; Hamann, C.; Buchner, B.; McCord, J.; Pelekhov, D. V .; Hammel, P. C. Imaging interfaces defined by abruptly varying internal magnetic fields by means of scanned nanoscale spin wav e modes. Phys. Rev. B 2015, 92, (21), 214413. (36) Lee, I.; Obukhov, Y .; Hauser, A. J.; Yang, F. Y .; Pelekhov, D. V .; Hammel, P. C. Nanoscale confined mode ferromagnetic resonance imaging of an individual Ni81Fe19 disk using magnetic resonance force micro scopy (invited). J. Appl. Phys. 2011, 109, (7), 07D313. (37) Song, Q. J.; Wang, H. F.; Xu, X. L.; Pan, X. C.; Wang, Y . L.; Song, F. Q.; Wan, X. G.; Dai, L. The polarization -dependent anisotropic Raman response of few -layer and bulk WTe2 under different excitation wavelengths. Rsc Adv 2016, 6, (105), 103830 -103837. (38) Tserkovnyak, Y .; Brataas, A.; Bauer, G. E. Enhanced Gilbert damping in thin ferromagnetic films. Phys. Rev. Lett. 2002, 88, (11), 117601. (39) Wu, G. Z.; Cheng, Y .; Guo, S. D.; Yang, F.; P elekhov, D.; Hammel, P. C. Nanoscale imaging of Gilbert damping using signal amplitude mapping. Appl. Phys. Lett. 2021, 118, (4), 042403. (40) Lee, A. J.; Ahmed, A. S.; Flores, J.; Guo, S. D.; Wang, B. B.; Bagues, N.; McComb, D. W.; Yang, F. Y . Probing th e Source of the Interfacial Dzyaloshinskii -Moriya Interaction Responsible for the Topological Hall Effect in Metal/Tm3Fe5O12 Systems. Phys. Rev. Lett. 2020, 124, (10), 107201. (41) Banerjee, S.; Rowland, J.; Erten, O.; Randeria, M. Enhanced Stability of S kyrmions in Two - Dimensional Chiral Magnets with Rashba Spin -Orbit Coupling. Physical Review X 2014, 4, (3), 031045. (42) Wang, Y . J.; Liu, E. F.; Liu, H. M.; Pan, Y . M.; Zhang, L. Q.; Zeng, J. W.; Fu, Y . J.; Wang, M.; Xu, K.; Huang, Z.; Wang, Z. L.; Lu, H . Z.; Xing, D. Y .; Wang, B. G.; Wan, X. G.; Miao, F. Gate -tunable negative longitudinal magnetoresistance in the predicted type -II Weyl semimetal WTe2. Nat. Commun. 2016, 7, (1), 1 -6. (43) Wu, Y . Y .; Zhang, S. F.; Zhang, J. W.; Wang, W.; Zhu, Y . L.; Hu, J .; Yin, G.; Wong, K.; Fang, C.; Wan, C. H.; Han, X. F.; Shao, Q. M.; Taniguchi, T.; Watanabe, K.; Zang, J. D.; Mao, Z. Q.; Zhang, X. X.; Wang, K. L. Neel -type skyrmion in WTe2/Fe3GeTe2 van der Waals heterostructure. Nat. Commun. 2020, 11, (1), 1-6. 22 (44) W ang, D. Y .; Che, S.; Cao, G. X.; Lyu, R.; Watanabe, K.; Taniguchi, T.; Lau, C. N.; Bockrath, M. Quantum Hall Effect Measurement of Spin -Orbit Coupling Strengths in Ultraclean Bilayer Graphene/WSe2 Heterostructures. Nano Letters 2019, 19, (10), 7028 -7034. (45) Zhao, Y . F.; Liu, H. W.; Yan, J. Q.; An, W.; Liu, J.; Zhang, X.; Wang, H. C.; Liu, Y .; Jiang, H.; Li, Q.; Wang, Y .; Li, X. Z.; Mandrus, D.; Xie, X. C.; Pan, M. H.; Wang, J. Anisotropic magnetotransport and exotic longitudinal linear magnetoresistance in WTe2 crystals. Phys. Rev. B 2015, 92, (4), 041104. 23 Enhancing Perpendicular Magnetic Anisotropy in Garnet Ferrimagnet by Interfacing with Few-Layer WTe 2 Guanzhong Wu1*, Dongying Wang1, Nish chhal Verma1, Rahul Rao2, Yang Cheng1, Side Guo1, Guixin Cao1, Kenji Watanabe3, Takashi Taniguchi4, Chun Ning Lau1, Fengyuan Yang1, Mohit Randeria1, Marc Bockrath1, and P. Chris Hammel1 1.Department of Physics, The Ohio State University, Columbus, OH, 43210, USA 2.Materials and Manufacturing Directorate, Air Force Research Laboratory, Wright -Patterson Air Force Base, Dayton, OH, 45433, USA 3. Research Center for Functional Materials, National Institute for Materials Science, 1 -1 Namiki, Tsukuba 305 -0044, Japan 4. International Center for Materials Nanoarchitectonics , National Institute for Materials Science, 1-1 Namiki, Tsukuba 305 -0044, Japan *wu.2314@osu.edu 24 FMRFM line-scan across edge of WTe 2 with different thickness Fig. S1 a-d, FMRFM line-scans along the traces 1~4 indicated in Fig. 3b respectively. The gray shaded area in four figures are outlining the location of WTe2 flake . 𝐻r,1 and ∆𝐴 at each position are derived by fitting the 𝑛=1 LM to a Lorentzian line shape. The thickness of WTe2 flake at each location are measured by atomic force microscope. 25 FMRFM measurement on Sample 2 Fig. S 2 FMRFM measurement on sample 2. a, The optical picture of the YIG/WTe 2/hBN heterostructure. b, 2D mapping of the 𝑛=1 LM resonance field in the black dash line circled area. c, 2D mapping of the Gilbert damping extracted from 𝑛=1 LM resonance peak amplitude in the black dash line circled area. d, FMRFM line scan along the trace indicated by the solid black line in Fig. S 2a. A constant background is subtracted to show only the signal from the LMs resonance. e, Fine scan zoomed in o n the quadruple layer WTe 2 stripe area 26 FMRFMR measurement at 4 GHz Fig. S 3 FMRFM measurement across WTe 2 edge at 4 GHz. FMRFM line scan is measured at 4 GHz across the WTe 2 flake edge. The shift of the resonance field 𝐻r,1 is 36 Oe, similar to the 𝐻r,1 shift measured at 2 GHz. This result excludes the possibility that the resonance field shift arises from modification of the gyromagnetic ratio. 27 Polarized Raman measurement Fig. S 4 Polarized Raman measurement. As shown by the red curve, the Raman spectrum taken on GGG/YIG/WTe 2/hBN heterostructure contains more peaks than WTe 2. The Raman spectrum taken in the GGG/YIG/hBN area identifies the peaks arising from the substrate GGG/YIG or top hBN encapsulation layer. By subtracting the Raman spectrum in the GGG/YIG/hBN area, the Raman spectra from WTe 2 layer are extracted and plotted in Fig. 2d. The black dash line are the markers indicating the Raman peaks of WTe 2 28 Impact of broken mirror reflection symmetr ies on the magnetic anisotropy We describe theo retical constraints on the interface -induced magnetic anisotropy in the WTe 2/YIG bilayer. We first show that symmetry arguments alone do not provide strong constraints on the anisotropy tensor, given that we are dealing with an interface between two crystalline materials at an arbitrary orientation with respect to each other . We then present qualitative arguments, based on the interfacia l spin - orbit coupling, that give insight into the magnetic anisotropy in WTe 2/YIG. This helps us understand why the easy-axis anisotropy that we observe in WTe 2/YIG differ s from the results of Lee et al. [1] on YIG interfaces with a dozen different metallic and semiconducting materials , all of which exhibit interface - induced easy-plane anisotropy as predicted by theory [2]. On general grounds, the anisotropy (free) energy can be written as ℱ𝑎𝑛𝑖𝑠= ∑ 𝐾𝑎𝑏𝑎,𝑏 𝑚𝑎 𝑚𝑏, (S1) where a and b take on values x,y,z. We focus here on the leading term, quadratic in the magnetization, and ignore higher order anisotropy terms like (mx4+my4+mz4) or (mx2my2+my2mz2+mz 2mx2). The form of 𝐾𝑎𝑏= 𝐾𝑏𝑎 is constrained by symmetry. Let us consider three cases , going from the most symmetric to the least . Case I: The only broken symmetry is interfacial inversion (z → - z), which is relevant for the experiments of Ref. [1]. The magnetization is an axial vector (or pseudovector) that transforms under rotation s like a vector but is unchanged under inversion . Thus (𝑚𝑥 ,𝑚𝑦 ,𝑚𝑧)→ (𝑚𝑥 ,− 𝑚𝑦 ,−𝑚𝑧) under reflection in a mirror plane with normal 𝑥̂. Using reflection symmetry in mirror planes normal to 𝑥̂ and to 29 𝑦̂ , we can see that all off -diagonal components of 𝐾𝑎𝑏 vanish. Further, f our-fold rotational symmetry about the 𝑧̂ axis shows that 𝐾𝑥𝑥= 𝐾𝑦𝑦. Using mx2+my2+mz2=1, we write 𝐾𝑥𝑥(m𝐱2+m𝐲2) in terms of m𝑧2, and d efining 𝐾𝑢= (𝐾𝑥𝑥− 𝐾𝑧𝑧), we obtain ℱ𝑎𝑛𝑖𝑠= − 𝐾𝑢 m𝑧2. (S2) This symmetry analysis only constrains the form of the anisotropy energy, but not the sign of 𝐾𝑢. We will give below a simple microscopic argument [2] that shows that 𝐾𝑢<0 (easy plane) for Case I. Case II: In addition to broken interfacial inversion (z → - z), let u s also break reflection symmetry in the plane normal to 𝑥̂. This would be the case if the crystalline axes of WTe 2 were aligned with YIG. This also breaks four-fold rotational symmetry about 𝑧̂, so that 𝐾𝑥𝑥 ≠ 𝐾𝑦𝑦. However, we can still use reflection symmetry in the plane normal to 𝑦̂ to conclude that 𝐾𝑥𝑦= 𝐾𝑦𝑧=0. Thus we find that K = (𝐾𝑥𝑥0 𝐾𝑥𝑧 0 𝐾𝑦𝑦0 𝐾𝑥𝑧0 𝐾𝑧𝑧) (S3) Case III: When the crystalline axes of WTe 2 are not aligned with YIG, which is the experimentally relevant case, all mirror reflection and rotation symmetries are broken. Then there are no symmetr y constrain ts on 𝐾𝑎𝑏 and all six components of this symmetric tensor are in general non -zero. Let us now see how, despite the lack of general symmetry -based constrai nts, we can still get some qualitative insight about the form of the anisotropy from simple microscopic considerations informed by symmetry. YIG is a ferrimagnetic Mott insulator, with two inequivalent Fe -sites coupled via 30 antiferromagnetic (AFM) superexch ange interactions. We thus focus on how interfacial spin -orbit coupling (SOC) impacts AFM superexchange. The broken symmetry at the interface leads to an electric field ℇ=−𝛁𝑉(𝒓), whose direction will be discussed in detail below for three cases. This in turn produces a magnetic field in the rest frame of the electron which underlies SOC. As the electron moves along 𝐫̂ij from site i to j, it experiences an SOC field in the direction 𝐝̂ij which is determined by ℇ ×𝐫̂ij . The SOC Hamiltonian is thus given by −𝑖𝜆∑𝑐𝐢𝛼†(𝐝̂ij∙𝝈𝛼𝛽)𝑐𝐣𝛃 𝛼𝛽 . Including the effect of this term in addition to the usual hopping t and Hubbard U in the standard strong coupling expansion calculation leads to the Hamiltonian ℋex=J∑𝐒i∙𝐒j <𝐢,𝐣>+D∑𝐝̂ij∙𝐒i×𝐒j <𝐢,𝐣> +K0∑(𝐝̂ij∙𝐒i)(𝐝̂ij∙𝐒j). <𝐢,𝐣> (S4) Here the spin 𝐒i at site i is coupled to its neighbors via the AFM superexchange 𝐽 ~𝑡2 𝑈 and the Dzyaloshinskii -Moriya interaction (DMI) 𝐷 ~𝑡𝜆 𝑈. The K0 term will be the focus of our attention belo w as it leads to magnetic anisotropy. We note that the general form of ℋex is in fact substantially independent [2] of the microscopic mechanism and very similar results are obtained not only for superexchange but also for Zener double exchange and RKKY interactions. Case I: Let us again return to the simplest case with broken interfacial inversion (z → - z). This leads to an electric field ℇ=−𝛁𝑉(𝒓) along ẑ , the normal to the interface. The SOC magnetic field direction is then given by 𝐝̂ij= ẑ ×𝐫̂ij ; see Figure S4(a). This is the well -known Rashba SOC at interfaces . We note in passing that 𝐝̂ij is antis ymmetric under the interchange of i and j, and thus leads to a DMI term where 𝐒i×𝐒j is also antisymmetric. 31 Fig. S 5 Symmetry based selection of magneto -crystalline anisotropy. Interfacial SOC originates from an effective Electric field ℇ=−𝛁𝑉(𝒓) whose direction is determined by the broken mirror planes in the system. This electric field leads to spin-orbit coupling (SOC), with the 𝐝̂ij= ℇ̂ ×𝐫̂ij, the direction of the SOC magnetic field. Note that the direction of the electron hop 𝐫̂ij lies in the xy plane of the interface. As shown in the text 𝐝̂ij controls the interface -induced magnetic anisotropy. (a) When only surface inversion is broken, 𝐝𝐢𝐣 is constrained to lie in the int erface and interfacial SOC leads to easy -plane anisotropy. (b) If there are other broken mirror planes, the 𝐝𝐢𝐣 must lie outside the interfacial plane. This can lead to a perpendicular magnetic anisotropy in systems like YIG/WTe 2 bilayers. We see that in Case I, 𝐝̂ij lies in the plane of the interface, and the third term in eq. (S 4) then takes the form K0∑(S𝐢𝑥𝑆𝐢+𝑦𝑥+S𝐢𝑦𝑆𝐢+𝑥𝑦) 𝐢 for a square lattice . To make the connection with magnetic anisotropy, we look at a continuum approximation with a s lowly varying magnetization 𝐦(𝐫). We make a Taylor expan sion of 𝐒r in terms of its value at 𝒓, denoted by 𝐦(𝐫), and its spatial derivatives . The exchange and DMI terms involve gradients of 𝐦(𝐫), but we focus here on local terms that do not involve derivatives to 32 understand the magnetic anisotropy . The leading term is + K0(m𝐱2+m𝐲2) which can be rewritten as – K0 m𝑧2 using the fact that mx2+my2+mz2=1 at each 𝒓. Thus, we may identify K0 with the anisotropy K𝑢 defined in eq. (S2). The microscopic analysis leads to the result K0= − 𝜆2 𝑈 < 0 and this explains the easy-plane anisotropy arising Rashba SOC at the interface . The easy-plane nature of the anisotropy is in fact a general feature of various microscopic models as emphasized in Ref. [2]. We note however that these author s use d the opposite sign convention for anisotropies from the one we use here . The easy plane vs. easy -axis character is , of course, independent of sign conventions. The FMR experiments of Ref. [1] have seen the interface -induced easy-plane anisotropy predicted by the theory in a YIG interfaces with several metallic and semiconducting materials . The key difference between the YIG/WTe 2 bilayer studied here and systems studied earlier [1] is that WTe 2 has a broken mirror plane (the ac plane ) as shown in Fig. 1(a) of the paper . We now look at the effect of this lower symmetry on the microscopic analysis. Case II: Let us break reflection symmetry in the plane normal to 𝑥̂ in addition to broken interfacial inversion. We choose x̂ parallel to the b axis, ŷ parallel to a, and ẑ parallel to c. Reflection symmetry in the ŷ mirror plane constrains the electric field ℇ =−𝛁𝑉(𝒓) to lie in the xz plane, at an angle 𝜃 from the z-axis as shown in Fig. S 5(b). Thus 𝐝ij=(sin𝜃𝑥̂+cos𝜃𝑧̂)×𝐫̂ij (S5) 33 where 𝐫̂ij is a vector in the interface (xy plane ) and 0≤𝜃≤𝜋. Using eq. (S5), we may rewrite the last term in the Hamiltonian (S4) as K0 sin2𝜃∑(S𝐢𝑧𝑆𝐢+𝑦𝑧) 𝐢+K0 cos2𝜃∑(S𝐢𝑥𝑆𝐢+𝑦𝑥+S𝐢𝑦𝑆𝐢+𝑥𝑦) 𝐢 −K0sin𝜃cos𝜃∑(S𝐢𝑧𝑆𝐢+𝑦𝑥+S𝐢𝑥𝑆𝐢+𝑦𝑧) 𝐢 As before, we make a continuum approximation with a smoothly varying 𝐦(𝐫) and focus only on the local terms, without gradients, to obtain the magnetic anisotropy . We find that the leading order contribution to anisotropy is −K0cos2𝜃m𝑧2+K0sin2𝜃 mzmx. This analysis correctly captures the non - zero K𝑥𝑧 expected on general grounds; see eq. (S3). We did not include here , for simplicity, the effects of broken four -fold rotation that would have led to 𝐾𝑥𝑥 ≠ 𝐾𝑦𝑦. Case III: When we lose all mirror symmetries, the case relevant to the YIG/WTe 2 experiment, the electric field ℇ =−𝛁𝑉(𝒓) will point in a general direction specified by 0≤𝜃≤𝜋 and 0≤𝜑≤2𝜋, and there will be no symmetry constraints on the anisotropy tensor 𝐾𝑎𝑏. Let us conclude by highlighting the key qualitative difference between Case I on the one hand and Cases II and III on the other . In Case I, the only broken symmetry is interfaci al inversion (z → - z). Then symmetry constrains the 𝐝̂ij, the direction of the SOC B-field, to lie in the plane of the interface and this leads to easy -plane anisotropy as described above. In Cases II and III, there are other additional broken mirror planes, and this leads to the 𝐝̂ij vector being pulled out of the plane of the interface. This immediately leads to the possibility of an easy -axis like character to the anisotropy, although in the general case one has a non -trivial anisotropy tensor 𝐾𝑎𝑏. 34 Reference [1] Lee, A. J.; Ahmed, A. S.; McCullian, B. A.; Guo, S. D.; Zhu, M. L.; Yu, S. S.; Woodward, P. M.; Hwang, J.; Hammel, P. C.; Yang, F. Y . Interfacial Rashba -Effect-Induced Anisotropy in Nonmagnetic - Material -Ferrimagnetic -Insulator Bilayers. Phys. Rev. Lett. 2020, 124, (25), 257202. [2] Banerjee, S.; Rowland, J.; Erten, O.; Randeria, M. Enhanced Stability of Skyrmions in Two - Dimensional Chiral Magnets with Rashba Spin -Orbit Coupling. Physical Review X 2014, 4, (3), 031045. | 2022-02-06 | Engineering magnetic anisotropy in a ferro- or ferrimagnetic (FM) thin film
is crucial in spintronic device. One way to modify the magnetic anisotropy is
through the surface of the FM thin film. Here, we report the emergence of a
perpendicular magnetic anisotropy (PMA) induced by interfacial interactions in
a heterostructure comprised of a garnet ferrimagnet, Y3Fe5O12 (YIG), and the
low-symmetry, high spin orbit coupling (SOC) transition metal dichalcogenide,
WTe2. At the same time, we also observed an enhancement in Gilbert damping in
the WTe2 covered YIG area. Both the magnitude of interface-induced PMA and the
Gilbert damping enhancement have no observable WTe2 thickness dependence down
to single quadruple-layer, indicating that the interfacial interaction plays a
critical role. The ability of WTe2 to enhance the PMA in FM thin film, combined
with its previously reported capability to generate out-of-plane damping like
spin torque, makes it desirable for magnetic memory applications. | Enhancing Perpendicular Magnetic Anisotropy in Garnet Ferrimagnet by Interfacing with Few-Layer WTe2 | 2202.02834v1 |
arXiv:1906.08987v1 [math.AP] 21 Jun 2019UNIQUE DETERMINATION OF THE DAMPING COEFFICIENT IN THE WAVE EQUATION USING POINT SOURCE AND RECEIVER DATA MANMOHAN VASHISTH Abstract. Inthisarticle, weconsidertheinverseproblemsofdetermi ningthedampingcoefficientappearing in the wave equation. We prove the unique determination of th e coefficient from the data coming from a single coincident source-receiver pair. Since our problem is under-determined, so some extra assumption on the coefficient is required to prove the uniqueness. Keywords : Inverse problems, wave equation, point source-receiver, d amping coefficient Mathematics subject classification 2010: 35L05, 35L10, 35R30, 74J25 1.Introduction We consider the following initial value problem (IVP), (/square−q(x)∂t)u(x,t) =δ(x,t) (x,t)∈R3×R u(x,t)|t<0= 0 x∈R3(1) where/square:=∂2 t−∆xdenotes the wave operator and the coefficient q∈C∞(R3) is known as damping coefficient. In this paper, we study the problem of determinat ion of coefficient qappearing in (1) from the knowledge of solution measured at a single point for a cer tain period of time. We are interested in the uniqueness of determination of coefficients qfrom the knowledge of u(0,t) fort∈[0,T] withT >0 in Equation (1). The problem studied here is motivated by geoph ysics, where geophysicists wish to determine the properties of earth structure by sending the waves from t he surface of the earth and measuring the correspondingscattered responses(see [2, 24] andreferen ces therein). Sincethecoefficient tobedetermined here depends on three variables while the given data depends on one variable as far as the parameter count is concerned, the problem studied here is under-determined . Thus some extra assumptions on coefficient q are required in order to make the inverse problem solvable. W e prove the uniqueness result for the radial coefficient. There are several results related to the inverse problems fo r the wave equation with point source. We list them here. Romanov in [18] considered the problem for determ ining the damping and potential coefficient in the wave equation with point source and proved unique dete rmination of these coefficients by measuring the solution on a set containing infinite points. In [12] the p roblem of determining the radial potential from the knowledge of solution measured on a unit sphere for s ome time interval is studied. Rakesh and Sacks in [16] established the uniqueness for angular contro lled potential in the wave equation from the knowledge of solution and its radial derivative measured on a unit sphere. In the above mentioned works the measurement set is an infinite set. Next we mention the wor k where uniqueness is established from the measurement of solution at a single point. Determinatio n of the potential from the data coming from a single coincident source-receiver pair is considered in [ 15] and the uniqueness result is established for the potentials which are either radial with respect a point differ ent from source location or the potentials which are comparable. Recently author in [25] extended the result of [15] to a separated point source and receiver data. To the best of our understanding, very few results exis t in the literature involving the recovery of the damping coefficient from point source and receiver data. O ur result, Theorem 1.1, is work in this direction. In the 1-dimensional inverse problems context, several results exist involving the uniqueness of 12 VASHISTH recovery of the coefficient which depends on the space variabl e corresponding to the first order derivative; see [9, 10, 11, 13, 19, 22]. We refer to [1, 3, 8, 14, 17] and refe rences therein for more works related to the point source inverse problems for the wave equation. We now state the main results of this article. Theorem 1.1. Suppose qi(x)∈C∞(R3),i= 1,2withqi(x) =Ai(|x|)for some C∞function Aion[0,∞). Letuibe the solution of the IVP (/square−qi(x)∂t)ui(x,t) =δ(x,t) (x,t)∈R3×R ui(x,t)|t<0= 0 x∈R3.(2) Ifu1(0,t) =u2(0,t)for allt∈[0,T]for some T >0, thenq1(x) =q2(x)for allxwith|x| ≤T/2, provided q1(0) =q2(0). The proof of the above theorem is based on an integral identit y derived using the solution to an adjoint problem as used in [21] and [23]. This idea was used in [4, 17, 2 5] as well. The article is organized as follows. In Section 2, we state th e existence and uniqueness results for the solution of Equation (1), the proof of which is given in [5, 8, 20]. Section 3contains the proof of Theorem 1.1. 2.Preliminaries Proposition 2.1. [5, pp.139,140] Suppose q∈C∞(R3)andu(x,t)satisfies the following initial value problem Pu(x,t) := (/square−q(x)∂t)u(x,t) =δ(x,t),(x,t)∈R3×R u(x,t)|t<0= 0, x ∈R3(3) thenu(x,t)is given by u(x,t) =R(x,t)δ(t−|x|) 4π|x|+v(x,t) (4) wherev(x,t) = 0fort <|x|and in the region t >|x|,v(x,t)is aC∞solution of the characteristic boundary value problem (Goursat Problem) Pv(x,t) = 0, for t > |x| v(x,|x|) =−R(x,|x|) 8π1/integraldisplay 0PR(sx,s|x|) R(sx,s|x|)ds,∀x∈R3(5) andR(x,t)is given by [5, pp. 134] R(x,t) = exp −1 21/integraldisplay 0q(sx)tds . (6) 3.Proof of Theorem 1.1 In this section, we prove Theorem 1.1. We will first prove an in tegral identity which will be used to prove our main result. Lemma3.1.Letui(x,t)fori= 1,2be the solution to Equation (2). Then the following integral identity holds for all σ≥0/integraldisplay R3/integraldisplay Rq(x)∂tu2(x,t)u1(x,2σ−t)dtdx=u(0,2σ) (7) whereq(x) :=q1(x)−q2(x)andu(x,t) = (u1−u2)(x,t).AN INVERSE PROBLEM WITH UNDER-DETERMINED DATA 3 Proof.Here we have usatisfies the following IVP /squareu(x,t)−q1(x)∂tu(x,t) =q(x)∂tu2(x,t) (x,t)∈R3×R u(x,t)|t<0= 0 x∈R3.(8) Multiplying Equation (8) by u1(x,2σ−t) and integrating over R3×R, we have /integraldisplay R3/integraldisplay Rq(x)∂tu2(x,t)u1(x,2σ−t)dtdx=/integraldisplay R3/integraldisplay R(/squareu(x,t)−q1(x)∂tu(x,t))u1(x,2σ−t)dtdx =/integraldisplay R3/integraldisplay Ru(x,t)(/squareu(x,t)−q1(x)∂tu1(x,2σ−t))dxdt where in the last step above we have used integration by parts and the properties of vin Proposition 2.1. Thus finally using the fact that u1is solution to (2), we get /integraldisplay R3/integraldisplay Rq(x)∂tu2(x,t)u1(x,2σ−t)dtdx=u(0,2σ); for all σ≥0. This completes the proof of the lemma. /square Using Lemma 3.1and the fact that u(0,t) = 0 for all t∈[0,T], we see that /integraldisplay R3/integraldisplay Rq(x)∂tu2(x,t)u1(x,2σ−t)dtdx= 0; for all σ∈[0,T/2]. Now using Equation (4), we get /integraldisplay R3/integraldisplay Rq(x)∂t/parenleftBigg R2(x,t)δ(t−|x|) 4π|x|+v2(x,t)/parenrightBigg/parenleftBigg R1(x,2σ−t)δ(2σ−t−|x|) 4π|x|/parenrightBigg dtdx +/integraldisplay R3/integraldisplay Rq(x)v1(x,2σ−t)∂t/parenleftBigg R2(x,t)δ(t−|x|) 4π|x|+v2(x,t)/parenrightBigg dtdx= 0.4 VASHISTH This gives /integraldisplay R3/integraldisplay Rq(x)∂tR2(x,t)R1(x,2σ−t)δ(t−|x|)δ(2σ−t−|x|) 16π2|x|2dtdx /bracehtipupleft /bracehtipdownright/bracehtipdownleft /bracehtipupright I1 +/integraldisplay R3/integraldisplay Rq(x)R2(x,t)R1(x,2σ−t)∂tδ(t−|x|)δ(2σ−t−|x|) 16π2|x|2dtdx /bracehtipupleft /bracehtipdownright/bracehtipdownleft /bracehtipupright I2 +/integraldisplay R3/integraldisplay Rq(x)∂t/parenleftBigR2(x,t)δ(t−|x|) 4π|x|/parenrightBig v1(x,2σ−t)dtdx /bracehtipupleft /bracehtipdownright/bracehtipdownleft /bracehtipupright I3 +/integraldisplay R3/integraldisplay Rq(x)∂tv2(x,t)R1(x,2σ−t)δ(2σ−t−|x|) 4π|x|dtdx /bracehtipupleft /bracehtipdownright/bracehtipdownleft /bracehtipupright I4 +/integraldisplay R3/integraldisplay Rq(x)∂tv2(x,t)v1(x,2σ−t)dtdx /bracehtipupleft /bracehtipdownright/bracehtipdownleft /bracehtipupright I5= 0; for all σ∈[0,T/2].(9) In a compact form, this can be written as I1+I2+I3+I4+I5= 0. (10) Next we simplify each Ijwithj= 1,2,.....,5. We will use the fact that vi(x,t) = 0 for t <|x|. We have I1=/integraldisplay R3/integraldisplay Rq(x)∂tR2(x,t)R1(x,2σ−t)δ(t−|x|)δ(2σ−t−|x|) 16π2|x|2dtdx =/integraldisplay |x|=σq(x)∂tR2(x,|x|)R1(x,|x|) 16π2|x|2dSx =−/integraldisplay |x|=σq(x)R1(x,|x|)R2(x,|x|) 32π2|x|2/parenleftBig1/integraldisplay 0q2(sx)ds/parenrightBig dSx. Next we simplify the integral I2. We use the following formula [7, Page 231, Eq.(10)] /integraldisplay δ′(r−|x|)ϕdx=−1 |x|2/integraldisplay |x|=r∂ ∂r/parenleftbig ϕr2/parenrightbig dSx. (11) Note that from this formula, by a change of variable, we have /integraldisplay δ′(2r−2|x|)ϕdx=−1 2|x|2/integraldisplay |x|=r∂ ∂r/parenleftbig ϕr2/parenrightbig dSx. (12)AN INVERSE PROBLEM WITH UNDER-DETERMINED DATA 5 Now I2=/integraldisplay R3/integraldisplay Rq(x)R2(x,t)R1(x,2σ−t)∂tδ(t−|x|)δ(2σ−t−|x|) 16π2|x|2dtdx =/integraldisplay R3/integraldisplay Rq(x)R2(x,t)R1(x,2σ−t)δ′(t−|x|)δ(2σ−t−|x|) 16π2|x|2dtdx =/integraldisplay R3q(x)R2(x,2σ−|x|)R1(x,|x|)δ′(2σ−2|x|) 16π2|x|2dx =−1 32π2σ2/integraldisplay |x|=σ∂ ∂r{q(x)R1(x,|x|)R2(x,2σ−|x|)}dSx. In the last step above, we used Equation (12). Next we have I3=/integraldisplay R3/integraldisplay Rq(x)∂t/parenleftBigR2(x,t)δ(t−|x|) 4π|x|/parenrightBig v1(x,2σ−t)dxdt. We can view the derivative above as a limit of the difference quo tients in the distribution topolgy [6, pp.48]. Combining this with the fact that v1isC2in{(x,t) :|x| ≤t}, we get, I3=−/integraldisplay R3/integraldisplay Rq(x)R2(x,t)δ(t−|x|) 4π|x|∂t/parenleftBig v1(x,2σ−t)/parenrightBig dxdt =/integraldisplay R3q(x)R2(x,|x|)∂tv1(x,2σ−|x|) 4π|x|dx. Again using the fact that v1(x,t) = 0 for t <|x|, we get, I3=/integraldisplay |x|≤σq(x)R2(x,|x|)∂tv1(x,2σ−|x|) 4π|x|dx. Next we simplify I4. Similiar to I3, we have I4=/integraldisplay R3/integraldisplay Rq(x)∂tv2(x,t)R1(x,2σ−t)δ(2σ−t−|x|) 4π|x|dtdx =/integraldisplay |x|≤σq(x)R1(x,|x|)∂tv2(x,2σ−|x|) 4π|x|dx. Finally, we have I5=/integraldisplay R3/integraldisplay Rq(x)∂tv2(x,t)v1(x,2σ−t)dtdx =/integraldisplay |x|≤σ2σ−|x|/integraldisplay |x|q(x)∂tv2(x,t)v1(x,2σ−t)dtdx.6 VASHISTH Now, we use the fact that qiis a radial function, that is, qi(x) =Ai(|x|). Then note that Ri(x,|x|) = exp −|x| 21/integraldisplay 0qi(sx)ds = exp −|x| 21/integraldisplay 0Ai(s|x|)ds is also radial. For simplicity, we denote R(x,|x|) byR(|x|). With this, we have I1=−A(σ)R1(σ)R2(σ) 8π1/integraldisplay 0A2(sσ)ds. Next we consider I2. First let us consider the derivative: Dr:=∂ ∂r(A(r)R1(x,r)R2(x,2σ−r)). After a routine calculation, we get, Dr=A′(r)R1(x,r)R2(x,r)−1 2A(r)2R1(x,r)R2(x,2σ−r) −σA(r)R1(x,r)R2(x,2σ−r)1/integraldisplay 0A′ 2(rs)sds =A′(r)R1(x,r)R2(x,r)−1 2A(r)2R1(x,r)R2(x,2σ−r) −A(r)R1(x,r)R2(x,2σ−r) σ r A2(r)−1/integraldisplay 0A2(rs)ds . On|x|=σ, we have Dr||x|=σ=R1(σ)R2(σ) A′(σ)−1 2A(σ)2−A(σ)A2(σ)+A(σ)1/integraldisplay 0A2(sσ)ds =R1(σ)R2(σ) A′(σ)−1 2A(σ)(A1+A2)(σ)+A(σ)1/integraldisplay 0A2(sσ)ds . Hence I2=−1 8π R1(σ)R2(σ) A′(σ)−1 2A(σ)(A1+A2)(σ)+A(σ)1/integraldisplay 0A2(sσ)ds . Let us denote ˜A(σ) =A(σ)R1(σ)R2(σ). Then I2=−1 8πd dσ˜A(σ)−1 8π˜A(σ)1/integraldisplay 0A2(sσ)ds. Therefore I1+I2=−1 8π 2˜A(σ)1/integraldisplay 0A2(sσ)ds+d dσ˜A(σ) .AN INVERSE PROBLEM WITH UNDER-DETERMINED DATA 7 Considering the following integrating factor for I1+I2 exp 2σ/integraldisplay 01/integraldisplay 0A2(ts)dtds , we have I1+I2=−1 8πexp −2σ/integraldisplay 01/integraldisplay 0A2(ts)dtds d dσ exp 2σ/integraldisplay 01/integraldisplay 0A2(ts)dtds ˜A(σ) . Now from Equation (10), we have 1 8πd dσ ˜A(σ)exp 2σ/integraldisplay 01/integraldisplay 0A2(st)dsdt = exp 2σ/integraldisplay 01/integraldisplay 0A2(st)dsdt /bracketleftBigg/integraldisplay |x|≤σq(x)R2(x,|x|)∂t{R1v1}(x,2σ−|x|) 4π|x|dx +/integraldisplay |x|≤σq(x)R1(x,|x|)∂tv2(x,2σ−|x|) 4π|x|dx +/integraldisplay |x|≤σ2σ−|x|/integraldisplay |x|q(x)∂tv2(x,t)v1(x,2σ−t)dtdx/bracketrightBigg for allσ∈[0,T/2].(13) Integrating on both sides with respect to σunder the assumption that ˜A(0) = 0, we get exp ˜σ/integraldisplay 01/integraldisplay 02A2(st)dsdt ˜A(˜σ) =˜σ/integraldisplay 0exp σ/integraldisplay 01/integraldisplay 02A2(st)dsdt /braceleftBigg/integraldisplay |x|≤σq(x)R2(x,|x|)∂tv1(x,2σ−|x|) 4π|x|dx +/integraldisplay |x|≤σq(x)R1(x,|x|)∂tv2(x,2σ−|x|) 4π|x|dx +/integraldisplay |x|≤σ2σ−|x|/integraldisplay |x|q(x)∂tv2(x,t)v1(x,2σ−t)dtdx/bracerightBigg dσ,for all ˜σ∈[0,T/2]. Now using the fact that R′ isare continuous, non-zero functions, and v′ isare continuous, we have the following inequality: |˜A(˜σ)| ≤C˜σ/integraldisplay 0|˜A(r)|drfor all ˜σ∈[0,T/2]. Now by Gronwall’s inequality, we have ˜A(σ) = 0 for all ˜ σ∈[0,T/2], which gives us q1(x) =q2(x) for all x∈R3such that |x| ≤T/2. This completes the proof.8 VASHISTH Acknowledgement The author would like to thank Dr. Venky Krishnan for useful d iscussions. He is supported by NSAF grant (No. U1530401). References [1] T. Aktosun, A. Machuca and P. Sacks; Determining the shap e of a human vocal tract from pressure measurements at the lips, Inverse Problems, vol. 33, (2017), 115002, 33 pages. [2] K.P. Bube and R. Burridge; The one-dimensional inverse p roblem of reflection seismology. SIAM Rev. 25 (1983), no. 4, 497–559. [3] R. Burridge; The Gel’fand-Levitan, the Marchenko, and t he Gopinath-Sondhi integral equations of inverse scatteri ng theory, regarded in the context of inverse impulse-respons e problems. Wave Motion 2 (1980), no. 4, 305–323. [4] E. Bl˚ asten; Well-posedness of the Goursat problem and s tability for point source inverse backscattering, 2017 Inv erse Problems 33 125003. [5] F.G. Friedlander; The wave equation on a curved space-ti me. Cambridge University Press, Cambridge-New York- Melbourne, 1975. Cambridge Monographs on Mathematical Phy sics, No. 2. [6] F.G. Friedlander and M. Joshi; The Theory of Distributio ns, 2nd Edition, Cambridge University Press, 1998. [7] I.M. Gel’fand and G. E. Shilov; Generalized functions. V ol. 1. Properties and operations. Translated from the 1958 R ussian original by Eugene Saletan. [8] M. M. Lavrent’ev, V. G. Romanov and S. P. Shishat ·ski˜i; Ill-posed problems of mathematical physics and analysis . Translated from the Russian by J. R. Schulenberger. [9] W. Ning and M. Yamamoto; The GelfandLevitan theory for on e-dimensional hyperbolic systems with impulsive inputs, Inverse Problems 24 (2008) 025004 (19pp). [10] Rakesh; An inverse impedance transmission problem for the wave equation. Comm. Partial Differential Equations 18 (1993), no. 3-4, 583–600. [11] Rakesh and P. Sacks; Impedance inversion from transmis sion data for the wave equation. Wave Motion 24 (1996), no. 3, 263–274. [12] Rakesh; Inversion of spherically symmetric potential s from boundary data for the wave equation. Inverse Problems 14 (1998), no. 4, 999–1007. [13] Rakesh; Characterization of transmission data for Web ster’s horn equation. Inverse Problems 16 (2000), no. 2, L9– L24. [14] Rakesh; An inverse problem for a layered medium with a po int source. Problems 19 (2003), no. 3, 497–506. [15] Rakesh; Inverse problems for the wave equation with a si ngle coincident source-receiver pair. Inverse Problems 24 (2008), no. 1, 015012, 16 pp. [16] Rakesh and P. Sacks; Uniqueness for a hyperbolic invers e problem with angular control on the coefficients. J. Inverse Ill-Posed Probl. 19 (2011), no. 1, 107–126. [17] Rakesh and G. Uhlmann; The point source inverse back-sc attering problem. Analysis, complex geometry, and mathema t- ical physics: in honor of Duong H. Phong, 279–289. [18] V.G. Romanov; On the problem of determining the coefficie nts in the lowest order terms of a hyperbolic equation. (Russian. Russian summary) Sibirsk. Mat. Zh. 33 (1992), no. 3, 156–160, 220; translation in Siberian Math. J. 33 (1992), no. 3, 497–500. [19] V. G. Romanov and D. I. Glushkova; The problem of determi ning two coefficients of a hyperbolic equation. (Russian) Dokl. Akad. Nauk 390 (2003), no. 4, 452–456. [20] V.G.Romanov; Integralgeometryandinverseproblemsf or hyperbolicequations, volume26.SpringerScienceandBu siness Media, 2013. [21] F. Santosa and W.W. Symes; High-frequency perturbatio nal analysis of the surface point-source response of a layer ed fluid. J. Comput. Phys. 74 (1988), no. 2, 318–381. [22] M.M. Sondhi; A survey of the vocal tract inverse problem : theory, computations and experiments. Inverse problems o f acoustic and elastic waves (Ithaca, N.Y., 1984), 1–19, SIAM , Philadelphia, PA, 1984. [23] P.D. Stefanov; A uniqueness result for the inverse back -scattering problem. Inverse Problems 6 (1990), no. 6, 1055 –1064. [24] W. W. Symes; The seismic reflection inverse problem. Inv erse Problems 25 (2009), no. 12, 123008, 39 pp. [25] M. Vashisth; An inverse problems for the wave equation w ith source and receiver at distinct points, Journal of Inver se and Ill-posed Problems, http://doi.org/10.1515/jiip-20 18-0004. Beijing Computational Science Research Center, Beijing 10 0193, China. E-mail: manmohanvashisth@gmail.com | 2019-06-21 | In this article, we consider the inverse problems of determining the damping
coefficient appearing in the wave equation. We prove the unique determination
of the coefficient from the data coming from a single coincident
source-receiver pair. Since our problem is under-determined, so some extra
assumption on the coefficient is required to prove the uniqueness. | Unique determination of the damping coefficient in the wave equation using point source and receiver data | 1906.08987v1 |
arXiv:2305.10111v1 [cond-mat.mes-hall] 17 May 2023Journal of the Physical Society of Japan FULL PAPERS Material Parameters for Faster Ballistic Switching of an In -plane Magnetized Nanomagnet Toshiki Yamaji1*and Hiroshi Imamura1 † 1National Institute of Advanced Industrial Science and Tech nology (AIST), Tsukuba, Ibaraki 305-8568, Japan High-speed magnetization switching of a nanomagnet is nece ssary for faster information processing. The ballistic switching by a pulsed magnetic filed is a promising candidate for the high-speed switching. It is known that the switch- ing speed of the ballistic switching can be increased by incr easing the magnitude of the pulsed magnetic field. However it is difficult to generate a strong and short magnetic field pulse in a sm all device. Here we explore another direction to achieve the high-speed ballistic switching by designing material parameters such as anisotropy constant, saturati on magnetization, and the Gilbert damping constant. We perfor m the macrospin simulations for the ballistic switching of in-plane magnetized nano magnets with varying material par ameters. The results are analyzed based on the switching dynamics on the energy density contour. We show that the puls e width required for the ballistic switching can be re- duced by increasing the magnetic anisotropy constant or by d ecreasing the saturation magnetization. We also show that there exists an optimal value of the Gilbert damping constan t that minimizes the pulse width required for the ballistic switching. 1. Introduction In modern information technologies huge amount of data are represented as the direction of the magnetization in a sm all magnet such as magnetic grains in magnetic tapes or hard disk drives. To write information on the conventional mag- netic recording media an external magnetic field is applied i n the opposite direction of the magnetization to switch the di - rection of the magnetization. During the switching the mag- netization undergoes multiple precessions around the loca l ef- fective field consisting of the external field, anisotropy fie ld, and demagnetizing field. The typical switching time or write time is of the order of nanoseconds. To meet the growing demand for fast information process- ing it is important to develop a faster switching scheme. The ballistic switching is a promising candidate for high-spee d switching, and much e ffort has been devoted to developing the ballistic switching both theoretically1–8)and experimen- tally.9–16)In ballistic switching a pulsed magnetic field is ap- plied perpendicular to the easy axis to induce the large-ang le precession around the external magnetic field axis. The dura - tion of the pulse is set to a half of the precession period. Aft er the pulse the magnetization relaxes to the equilibrium dire c- tion opposite to the initial direction. The switching speed of the ballistic switching can be increased by increasing the m ag- nitude of the pulsed field. However, it is di fficult to generate a strong and short field pulse in a small device. It is desired to find a way to speed up the ballistic switching without increas - ing magnetic field. The magnetization dynamics of the ballistic switching is determined by the torques due to the external magnetic field, the uniaxial anisotropy field, the demagnetizing field, and t he Gilbert damping. The torques other than the external mag- netic field torque are determined by the material parameters such as the anisotropy constant, the saturation magnetizat ion, and the Gilbert damping constant. There is room to speed up *toshiki-yamaji@aist.go.jp †h-imamura@aist.go.jpthe ballistic switching by designing the appropriate mater ial parameters. In the early 2000s the several groups each independently reported the optical microscope measurements of the ballis - tic switching by picosecond pulse magnetic field.9–13)Then the mechanism of a ballistic switching was analyzed in terms of the nonlinear dynamics concepts such as a fixed point, at- tractors, and saddle point.2, 3, 6)Especially the minimal field required for a ballistic switching was investigated by comp ar- ing the so-called Stoner-Wohlfarth (SW) type.2, 3)The damp- ing constant dependence of the minimal switching field was also studied.2)The characteristics of the parameters of a pulse magnetic field, i.e., magnitude, direction, and rise /fall time on the mechanism of a ballistic switching had been also studied by the simulations and experiments.6, 7, 14, 15) As described above, in 2000s and 2010s a ballistic switch- ing technique had received much attention for the fast magne - tization reversal with ringing suppression by fine-tuning t he magnetic pulse parameters. Due to the recent advance of an ultra-fast measurement17)the studies of a ballistic switching have attracted much attention again. Last year the in-plane magnetization switching dynamics as functions of the pulse magnetic field duration and amplitude was calculated and analyzed by using the conventional Landau-Lifshitz-Gilbe rt (LLG) equation and its inertial form, the so-called iLLG equation.16)The solutions of both equations were compared in terms of the switching characteristics, speed and energy density analysis. Both equations return qualitatively sim ilar switching dynamics. However the extensive material param- eter dependences of a ballistic switching region have not yet been sufficiently explored. Therefore it is worth clearing the extensive material parameter dependences of the ballis tic switching of an in-plane magnetized nanomagnet. In this paper, we study the ballistic switching of an in- plane magnetized nanomagnet with systematically varying the material parameters by using the macrospin simulations . The results show that the pulse width required for the bal- listic switching can be reduced by increasing the magnetic 1J. Phys. Soc. Jpn. FULL PAPERS Hp mz yx(a) (c) my at t = 10 ns (b) (d) 0 200 400-1 01 t [ps]my tp [ps]0 1 2 3 4 5tSW [ps] 110 10 210 3 tl tutSW 0 1 -1 0 1 2 3 4 502.55.010.0 7.5 tp [ps]Hp [T] Fig. 1. (a) Schematic illustration of the in-plane magnetized nano magnet. The pulse field, Hp, is applied along the x-direction. The initial direction of the magnetization is in the positive y-direction. (b) Gray scale map of myat t=10 ns as a function of the pulse field width, tp, and Hp. The black and white regions represent the success and failure of switchin g. The parameters areµ0Ms=0.92 T,µ0HK=0.1 T, andα=0.023. (c) Typical example of the time evolution of mywhen the magnetization switches ( Hp=5 T and tp =0.4 ps). The switching time, tSW, is defined as the time when mychanges the sign. (d) tpdependence of tSWalong the dashed horizontal line at Hp=5 T shown in Fig. 1(b). tlandtuare 3.15 ps and 3.93 ps, respectively. tSWat tl≤tp≤tuis 1.7 ps. anisotropy constant or by decreasing the saturation magnet i- zation. There exists an optimal value of the Gilbert damping constant that minimizes the pulse width required for ballis - tic switching. The simulation results are intuitively expl ained by analyzing the switching trajectory on the energy density contour. 2. Model and Method In this section we show the theoretical model, the numer- ical simulation method, and the analysis using the trajecto ry in the limit ofα→0. The macrospin model of the in-plane magnetized noanomagnet and the equations we solve to simu- late the magnetization dynamics are given in Sec. 2.1. In Sec . 2.2 we show that the switching conditions can be analyzed by using the trajectory on the energy density contour in the lim it ofα→0 if theα≪1. 2.1 Macrospin Model Simulation Figure 1(a) shows the schematic illustration of the in- plane magnetized nanomagnet. The pulsed magnetic field, Hp, is applied along the x-direction. The unit vector m= (mx,my,mz) indicates the direction of the magnetization. The size of the nanomagnet is assumed to be so small that the dy- namics of mcan be described by the macrospin LLG equation dm dt=−γm×/parenleftBigg Heff−α γdm dt/parenrightBigg , (1) where tis time,γis the gyromagnetic ratio, αis the Gilbert damping constant. The e ffective field, Heff=Hp+HK+Hd, comprises the pulse field, Hp, the anisotropy field, HK, andthe demagnetizing field, Hd. The anisotropy field and the de- magnetizing field are defined as HK=/bracketleftbig2K/(µ0Ms)/bracketrightbigmyey, (2) and Hd=µ0Msmzez, (3) respectively, where Kis the uniaxial anisotropy constant, µ0 is the magnetic permeability of vacuum, Msis the saturation magnetization, and ejis the unit vector along the j-axis ( j= x,y,z). The switching dynamics are calculated by numerically solving the LLG equation. The initial ( t=0) direction is set asmy=1. The rectangular shaped pulse magnetic field with duration of tpis applied at t=0. The time evolution of magne- tization dynamics are calculated for 10 ns. Success or failu re of switching is determined by whether my<−0.5 att=10 ns. Figure 1(b) shows the gray scale plot of myatt=10 ns on the tp-Hpplane. Following Ref. 16 the parameters are as- sumed to beµ0Ms=0.92 T, K=2.3 kJ/m3, i.e.µ0HK= 0.1 T, andα=0.023. The black and white regions represent the success and failure of switching, respectively. The wid e black region at upper right of Fig. 1(b) represents the balli stic switching region (BSR). A typical example of the time evolu- tion of mywhen the magnetization switches is shown in Fig. 1(c). The switching time, tSW, is defined as the time when my changes the sign. Figure 1(d) shows the tpdependence of tSW along the horizontal line shown in Fig. 1(b), i.e. at Hp=5 T. The BSR indicated by shade appears between tl=3.15 ps and tu=3.93 ps, where tSW=1.7 ps independent of tp. The lower and upper boundary of the BSR are represented by tlandtu, respectively. We investigate the material parameter dependence of tlandtuwith keeping Hp=5 T. 2.2 Analysis of the Switching Conditions for α≪1 If the Gilbert damping constant is much smaller than unity the approximate value of tlandtucan be obtained without performing macrospin simulations. In the limit of α→0, the trajectory is represented by the energy contour because the en- ergy is conserved during the motion of m. The energy density, E, of the nanomagnet is defined as18) E=1 2µ0M2 scos2θ+K(1−sin2θsin2φ), (4) whereθandφare the polar and azimuthal angles of the mag- netization, respectively. The color plot of the energy dens ity contour is shown in Fig. 2. The separatrix representing the energy contour with E=Kis indicated by the white curve, which is expresses as 1 2µ0M2 scos2θ−Ksin2θsin2φ=0. (5) The green dot indicates the initial direction of matt=0. The black curve represents the trajectory of munder the pulse field ofHpin the limit ofα→0. Under the pulse field the energy density is given by E=1 2µ0M2 scos2θ+K(1−sin2θsin2φ) −µ0MsHpsinθcosφ. (6) 2J. Phys. Soc. Jpn. FULL PAPERS 01 5 4 3 26E/K tltu θ φ Fig. 2. (Color online) Color plot of the energy density contour give n by Eq. (4).θandφare the polar and azimuthal angles of the magnetization, re- spectively. The material parameters, MsandKare same as in Fig. 1. The separatrix given by Eq. (5) is indicated by the white curve. T he initial direc- tion of mis indicated by the green dot at ( θ,φ)=(π/2,π/2). The black curve represents the trajectory of the magnetization under the fie ld of Hp=5 T in the limit ofα→0, which is given by Eq. (7). The yellow stars indicate the intersection points of the separatrix and the trajectory, w hich correspond to tp =tlandtu. If the pulse is turned o ffattl≤t≤tu, the magnetization switches ballistically. The yellow triangle indicates the turning p oint of the trajectory of the magnetization near mz=1, at whichφ=0. Since the energy density of the initial direction, θ=φ=π/2, isE=0, the trajectory under the pulse field is expressed as 1 2µ0M2 scos2θ+K(1−sin2θsin2φ) −µ0MsHpsinθcosφ=0. (7) The yellow stars indicate the points where the trajectory crosses the separatrix surrounding the equilibrium point a t φ=−π/2. The upper and lower points indicates the direc- tion of mat the end of the pulse with tp=tuandtl, re- spectively. The corresponding angles ( θl,φl) and (θu,φu) can be obtained by solving Eqs. (5) and (7) simultaneously. If tl≤tp≤tu, the magnetization relaxes to the equilibrium di- rection at (θ,φ)=(π/2,−π/2) after the pulse to complete the switching. We can obtain the approximate expressions of tl andtuas follows. Assuming that the pulse field is much larger than the other fields, the angular velocity of the precession ,ω, is approximated as γHp/(1+α2), and tlandtuare analytically obtained as tl=π−2θturn ω−1 2∆θ ω, (8) and tu=π−2θturn ω+1 2∆θ ω, (9) where∆θ=θu−θl, andθturnis the polar angle at the turning point (φ=0) indicated by the yellow triangle.3. Results and Discussion In this section we discuss the dependence of the BSR on the material parameters by analyzing the numerical simula- tion results and Eqs. (8) and (9). The results for the variati on of the magnetic anisotropy constant, K, saturation magnetiza- tion, Ms, and the Gilbert damping constant, α, will be given in Secs. 3.1, 3.2, and 3.3, respectively. 3.1 Anisotropy Constant Dependence of the BSR Figure 3(a) shows the anisotropy constant, K, dependence of the BSR. The parameters are Hp=5 T,µ0Ms=0.92 T, and α=0.023. The simulation results of tlandtuare indicated by the orange and blue dots, respectively. The analytical ap - proximations of tlandtuobtained by solving Eqs. (5),(7),(8), and (9) are represented by the orange and blue curves, respec - tively. The simulation and analytical results agree well wi th each other because the Gilbert damping constant is as small a s 0.023. As shown in Fig. 3(a), tlis a monotonically decreasing function of Kwhile tuis a monotonically increasing function ofK. As a result the width of the BSR, tu-tl, is a monoton- ically increasing function of Kas shown in the inset of Fig. 3(a). In the left panel of Fig. 3(b) the separatrix and the trajecto ry withα=0 for K=2.3 kJ/m3are shown by the blue and black curves, respectively. The same plot for K=9.3 kJ/m3 is shown in the right panel. As shown in these panels, the increase of Kdoes not change the trajectory much. However, the increase of Kchanges the separatrix significantly through the second term of Eq. (5). Assuming that the angular velocit y of the precession is almost constant, the spread of the area surrounded by the separatrix results in the spread of the tim e difference between tlandtu. As a result the BSR is spread by the increase of Kas shown in Fig. 3(a) 3.2 Saturation Magnetization Dependence of the BSR Figure 4(a) shows the saturation magnetization dependence of the BSR obtained by the numerical simulation and the ana- lytical approximations. The horizontal axis represents th e sat- uration magnetization in unit of T, i.e µ0Ms. The parameters areHp=5 T,K=2.3 kJ/m3, andα=0.023. The symbols are the same as in Fig. 3(a). The lower boundary of the BSR, tl, increases as theµ0Msincreases while the upper boundary of the BSR, tu, decreases with increase of µ0Ms. Therefore, the faster switching is available for smaller Ms. Theµ0Msdepen- dence of the BSR ( tu-tl) is also shown in the inset of Fig. 4(a). The BSR decreases with increase of µ0Ms. In other words, the wider BSR is obtained for smaller Ms. In the right panel of Fig. 4(b) the separatrix and the trajec- tory withα=0 forµ0Ms=0.35 T are shown by the blue and black curves, respectively. The same plot for µ0Ms=0.92 T is shown in the left panel. As shown in these panels, the increas e ofMsdoes not change the trajectory much but decrease the separatrix significantly through the first term of Eq. (5). As - suming that the angular velocity of the precession is almost constant, the reduction of the area surrounded by the separa - trix results in the reduction of the time di fference between tl andtu. As a result the BSR decreases with increase of Msas shown in Fig. 4(a) 3J. Phys. Soc. Jpn. FULL PAPERStl, t u [ps] 0 10 20 30 40 2.03.04.05.0 K [kJ/m3]ballistic switching region (a) tltu tu - t l [ps] K [kJ/m 3]0 10 20 30 40 0.0 0.5 1.0 1.5 2.0 2.5 0π π/2 0 0 π π -π -π π /2 π/2 -π /2 -π /2 θ φ φˑˑ ˑˑ(b) K = 2.3 kJ/m3K = 9.3 kJ/m3 tltltutu Fig. 3. (Color online) (a) Anisotropy constant, K, dependence of the BSR (orange shade). Simulation results of tlandtuare plotted by the orange and blue dots, respectively. The analytical results are indica ted by the solid curves with the same color. The parameters are Hp=5 T,µ0Ms=0.92 T, andα= 0.023. In the inset the simulation and analytical results of the width of the BSR, tu-tl, are plotted by the dots and the solid curve, respectively. ( b) Typical examples of the trajectory of the magnetization (bl ack curve) and the separatrix (blue curve). The left and right panels show the r esults for K=2.3 kJ/m3andK=9.3 kJ/m3, respectively. The orange and blue stars indicate the direction at t=tlandtu, respectively. The green dots indicate the initial direction of m. 3.3 Gilbert Damping Constant Dependence of the BSR Figure 5(a) shows the simulation results of the Gilbert damping constant, α, dependence of the BSR. The width of the BSR is shown in the inset. The symbols are the same as in Fig. 3(a). The approximate values obtained by Eqs. (8) and (9) are not shown because the αis not limited toα≪1. The parameters are Hp=5 T, K=2.3 kJ/m3, andµ0Ms=0.92 T. There exists an optimal value of αthat minimizes tl. The optimum value in Fig. 5 (a) is αopt=0.35. To understand the mechanism for minimization of tlat a certain value ofαone need to consider two di fferent effects of αon the magnetization dynamics. One e ffect is the decrease of the precession angular velocity with increase of α. The pre- cession angular velocity around the e ffective field of Heffis given by (γHeff)/(1+α2), which decreases with increase of α. This effect causes the increase of tlandtu. The other effect is the increase of the energy dissipation rate with increase ofα. Let us consider the trajectory in the cases of small damping ( α=0.023) and large damping ( α=αopt). In Fig. 5 (b) the typical examples of the trajectory for the small damping are shown by the yellow and green curves and dots on the energy density contour. The pulse widths are tp=tl(=3.15 ps) and 3.14 ps. The trajectories during the pulse are represented by the solid curves and the trajectori es after the pulse are represented by the dots. The white curve shows the separatrix and the black dot indicates the initial di- tl, t u [ps] 2.03.04.05.0 0.0 0.3 0.6 0.9 1.2 μ0Ms [T]ballistic switching region tltu(a) (b) μ0Ms = 0.92 T μ0Ms = 0.35 T 0π π/2 0 0 π π -π -π π /2 π/2 -π /2 -π /2 θ φ φˑˑ ˑˑ tl tltu tu1.5tu - t l [ps] 0123 0.0 0.3 0.6 0.9 1.2 μ0Ms [T] 1.5 Fig. 4. (Color online) (a) Saturation magnetization dependence of the BSR. The horizontal axis represents the saturation magneti zation in unit of T, i.eµ0Ms. The parameters are Hp=5 T, K=2.3 kJ/m3, andα=0.023. The symbols are the same as in Fig. 3 (a). (b) Typical examples of t he trajectory of the magnetization (black curve) and the separatrix (blue curve). The right and left panels show the results for µ0Ms=0.35 T and 0.92 T, respectively. The symbols are the same as in Fig. 3 (b). rection. The yellow and green stars indicate the points wher e the trajectories cross the separatrix surrounding the targ et and initial states, respectively. The arrows indicate the dire ction of the movement of the magnetization. For the small damp- ing, even very close to the separatrix around the target stat e at the end of the pulse, the magnetization flows to the sepatrari x around the initial state and relax to the initial state after many precessions with the slow energy dissipation. Figure 5 (c) shows the tpdependence of tSWat the large damping (α=αopt). All parameters except αare the same as in Fig. 1 (d). t′ l,tl, and tuare 0.82 ps, 1.98 ps, and 4.54 ps, respectively. t′ lis the time when for the large damping the magnetization goes across the e ffective separatrix around the initial state during the pulse duration. In Fig. 5 (d) the typ ical examples of the trajectory for the large damping are shown by the yellow ( tp=0.9 ps), green ( tp=tl=1.98 ps), and purple ( tp=4.55 ps) curves and dots on the energy density contour. The symbols are the same as in Fig. 5 (b). In the region 1 ( tp<t′ l) of Fig. 5 (c) the magnetization is located on the inside of the effective separatrix at the end of the pulse and return to the initial state. The trajectory is not shown in Fi g. 5 (d). In the region 2 ( t′ l≤tp<tl) of Fig. 5 (c) after the pulse is removed the magnetization move toward the target state unde r the effective field of Heffand goes across the separatrix. Then the magnetization finishes the switching. The typical examp le of the trajectory is shown by the yellow curve and dots in Fig. 5 (d). In the region 3 ( tl≤tp≤tu) after the pulse is removed the magnetization ballistically switches. The typical exa mple 4J. Phys. Soc. Jpn. FULL PAPERS tl, t u [ps] 2.04.06.08.0 0.0 0.2 0.4 0.6 0.8 αballistic switching region tltu(a) 0 tu - t l [ps] 0.00123 0.2 0.4 0.6 0.8 α45 αopt αopt 01 5 4 3 26E/K α = 0.023 tp = t l = 3.15 ps tp = 3.14 ps (b) 0 -π/2 π/2 π -π0π/2 π φθ < << < < ˑ ˑ (c) tp [ps] 0 1 2 3 4 5tSW [ps] 110 10 210 3 tl tu1 2 4 3 tl’ 01 5 4 3 26E/K α = 0.35 tp = 0.9 ps tp = t l = 1.98 ps tp = 4.55 ps (d) 0 -π/2 π/2 π -π0π/2 π φθ < < <ˑˑ < < < ˑ Fig. 5. (Color online) (a) Simulation results of the tlandtuas a function ofα. The parameters are Hp=5.0 T, K=2.3 kJ/m3, andµ0Ms=0.92 T. The symbols are the same as in Fig. 3 (a). (b) The trajectori es atα= 0.023 with tp=3.15 ps (yellow) and 3.14 ps (green) on the energy density contour. The trajectory during the pulse is represented by t he solid curve. The trajectory after the pulse is represented by the dots. The wh ite curve shows the separatrix.The direction of the trajectory is indicate d by the arrow. The star indicates the intersection point of the trajectory and the separatrix. The initial direction is indicated by the black dot. (c) tpdependence of tSWat α=αopt. All parameters except αare the same as in Fig. 1 (d). t′ l,tl, and tu are 0.82 ps, 1.98 ps, and 4.54 ps, respectively. tSWattl≤tp≤tuis 1.98 ps. (d) The trajectories at α=αoptwith tp=0.9 ps (yellow), 1.98 ps (green), and 4.55 ps (purple) on the energy density contour. The symbols a re the same as in Fig. 5 (b). of the trajectory is shown by the green curve and dots in Fig.5 (d). As explained in Sec. 2.1, in this study the tSWis defined as the time when mychanges the sign, in other words, the magnetization reaches the turning point ( φ=0) on the energy density contour. Therefore the tlfor the large damping is equal to the ballistic tSW, i.e. the time when under the pulse field the magnetization reaches the turning point. The ballistic tSWat the region 3 is 1.98 ps. As described above, for the large damping the magnetiza- tion can relax to the target state even at greater distances f rom the separatrix by moving under Heffwith the fast energy dis- sipation. The effect can be regarded as the e ffective spread of the separatrix and results in the decrease of tland the in- crease of tuwith increase ofα. Therefore, there exists the optimal value ofαthat minimizes tlwhile tumonotonically increases with increase of αas shown in Fig. 5(a). In the re- gion 4 ( tp>tu) after the pulse is removed the magnetization moves toward the separatrix around the initial state under Heff and relaxes to the initial state. We find that the BSR for the large damping can be explained by the anisotropic spread of the effective separatrix with increasing α, which is fundamen- tally due to the breaking of the spatial inversion symmetry o f the spin dynamics. The broken symmetry of the spatial inver- sion of the spin dynamics for the large damping can be easily confirmed by comparing Fig. 5 (c) with Fig. 1 (d). 4. Summary In summary, we study the material parameter dependence of the ballistic switching region of the in-plane magnetize d nanomagnets based on the macrospin model. The results show that the pulse width required for the ballistic switching ca n be reduced by increasing the magnetic anisotropy constant or b y decreasing the saturation magnetization. The results also re- vealed that there exists an optimal value of the Gilbert damp - ing constant that minimizes the pulse width required for the ballistic switching. The simulation results are explained by analyzing the trajectories on the energy contour. The resul ts are useful for further development of the high-speed inform a- tion processing using the ballistic switching of magnetiza tion. This work is partially supported by JSPS KAKENHI Grant Number JP20K05313. 1) L. He and W. D. Doyle: J. Appl. Phys. 79(1996) 6489. 2) Z. Z. Sun and X. R. Wang: Phys. Rev. B 71(2005) 174430. 3) D. Xiao, M. Tsoi, and Q. Niu: Journal of Applied Physics 99(2006) 013903. 4) Y . Nozaki and K. Matsuyama: Jpn. J. Appl. Phys. 45(2006) L758. 5) Y . Nozaki and K. Matsuyama: Journal of Applied Physics 100(2006) 053911. 6) Q. F. Xiao, B. C. Choi, J. Rudge, Y . K. Hong, and G. Donohoe: J ournal of Applied Physics 101(2007) 024306. 7) P. P. Horley, V . R. Vieira, P. M. Gorley, V . K. Dugaev, J. Ber akdar, and J. Barna´ s: Journal of Magnetism and Magnetic Materials 322(2010) 1373. 8) Y . B. Bazaliy: Journal of Applied Physics 110(2011) 063920. 9) T. Gerrits, H. A. M. van den Berg, J. Hohlfeld, L. B¨ ar, and T . Rasing: Nature 418(2002) 509. 10) I. Tudosa, C. Stamm, A. B. Kashuba, F. King, H. C. Siegmann , J. St¨ ohr, G. Ju, B. Lu, and D. Weller: Nature 428(2004) 831. 11) H. W. Schumacher, C. Chappert, R. C. Sousa, P. P. Freitas, and J. Miltat: Phys. Rev. Lett. 90(2003) 017204. 12) W. K. Hiebert, L. Lagae, J. Das, J. Bekaert, R. Wirix-Spee tjens, and J. De Boeck: Journal of Applied Physics 93(2003) 6906. 13) W. K. Hiebert, L. Lagae, and J. De Boeck: Phys. Rev. B 68(2003) 5J. Phys. Soc. Jpn. FULL PAPERS 020402. 14) H. W. Schumacher: Appl. Phys. Lett. 87(2005) 042504. 15) N. Kikuchi, Y . Suyama, S. Okamoto, O. Kitakami, and T. Shi matsu: Appl. Phys. Lett. 104(2014) 112409. 16) K. Neeraj, M. Pancaldi, V . Scalera, S. Perna, M. d’Aquino , C. Serpico, and S. Bonetti: Phys. Rev. B 105(2022) 054415.17) K. Neeraj, N. Awari, S. Kovalev, D. Polley, N. Zhou Hagstr ¨ om, S. S. P. K. Arekapudi, A. Semisalova, K. Lenz, B. Green, J.-C. Dein ert, I. Ilyakov, M. Chen, M. Bawatna, V . Scalera, M. d’Aquino, C. S erpico, O. Hellwig, J.-E. Wegrowe, M. Gensch, and S. Bonetti: Nat. Ph ys.17 (2021) 245. 18) W. F. Brown: Phys. Rev. 130(1963) 1677. 6 | 2023-05-17 | High-speed magnetization switching of a nanomagnet is necessary for faster
information processing. The ballistic switching by a pulsed magnetic filed is a
promising candidate for the high-speed switching. It is known that the
switching speed of the ballistic switching can be increased by increasing the
magnitude of the pulsed magnetic field. However it is difficult to generate a
strong and short magnetic field pulse in a small device. Here we explore
another direction to achieve the high-speed ballistic switching by designing
material parameters such as anisotropy constant, saturation magnetization, and
the Gilbert damping constant. We perform the macrospin simulations for the
ballistic switching of in-plane magnetized nano magnets with varying material
parameters. The results are analyzed based on the switching dynamics on the
energy density contour. We show that the pulse width required for the ballistic
switching can be reduced by increasing the magnetic anisotropy constant or by
decreasing the saturation magnetization. We also show that there exists an
optimal value of the Gilbert damping constant that minimizes the pulse width
required for the ballistic switching. | Material Parameters for Faster Ballistic Switching of an In-plane Magnetized Nanomagnet | 2305.10111v1 |
Probing spin dynamics of ultra-thin van der Waals magnets via photon-magnon coupling Christoph W. Zollitsch,1,a)Safe Khan,1Vu Thanh Trung Nam,2Ivan A. Verzhbitskiy,2Dimitrios Sagkovits,1, 3 James O’Sullivan,1Oscar W. Kennedy,1Mara Strungaru,4Elton J. G. Santos,4, 5John J. L. Morton,1, 6Goki Eda,7, 2, 8and Hidekazu Kurebayashi1, 6, 9 1)London Centre for Nanotechnology, University College London, 17-19 Gordon Street, London, WCH1 0AH, UK 2)Department of Physics, Faculty of Science, National University of Singapore, 2 Science Drive 3, Singapore 117542, Singapore 3)National Physical Laboratory, Hampton Road, Teddington TW11 0LW, UK 4)Institute for Condensed Matter Physics and Complex Systems, School of Physics and Astronomy, The University of Edinburgh, Edinburgh EH9 3FD, UK 5)Higgs Centre for Theoretical Physics, The University of Edinburgh, Edinburgh EH9 3FD, UK 6)Department of Electronic & Electrical Engineering, UCL, London WC1E 7JE, United Kingdom 7)Centre for Advanced 2D Materials, National University of Singapore, 6 Science Drive 2, Singapore 117546, Singapore 8)Department of Chemistry, Faculty of Science, National University of Singapore, 3 Science Drive 3, Singapore 117543, Singapore 9)WPI Advanced Institute for Materials Research, Tohoku University, 2-1-1, Katahira, Sendai, 980- 8577, Japan (Dated: 1 May 2023) Layered van der Waals (vdW) magnets can maintain a magnetic order even down to the single-layer regime and hold promise for integrated spintronic devices. While the magnetic ground state of vdW magnets was extensively studied, key parameters of spin dynamics, like the Gilbert damping, crucial for designing ultra-fast spintronic de- vices, remains largely unexplored. Despite recent studies by optical excitation and detection, achieving spin wave control with microwaves is highly desirable, as modern in- tegrated information technologies predominantly are op- erated with these. The intrinsically small numbers of spins, however, poses a major challenge to this. Here, we present a hybrid approach to detect spin dynamics medi- ated by photon-magnon coupling between high-Q super- conducting resonators and ultra-thin flakes of Cr 2Ge2Te6 (CGT) as thin as 11 nm. We test and benchmark our tech- nique with 23 individual CGT flakes and extract an upper limit for the Gilbert damping parameter. These results are crucial in designing on-chip integrated circuits using vdW magnets and offer prospects for probing spin dynamics of monolayer vdW magnets. INTRODUCTION van der Waals (vdW) materials1–3consist of individual atomic layers bonded by vdW forces and can host different types of collective excitations such as plasmons, phonons and magnons. Strong coupling between these excitation modes and electromagnetic waves (i.e. photonic modes) creates con- fined light-matter hybrid modes, termed polaritons. Polaritons a)Electronic mail: c.zollitsch@ucl.ac.ukin vdW materials are an ideal model system to explore a va- riety of polaritonic states5,6, e.g. surface plasmon polaritons in graphene7,8and exciton polaritons in a monolayer MoS 2 embedded inside a dielectric microcavity9. These states can be further modified by electrostatic gating16, as well as by hetero-structuring with dissimilar vdW layers1. Numerous studies on magnon polaritons (MPs)11,12have been using macroscopic yttrium iron garnet (YIG) cou- pled to either three-dimensional cavities13or to on-chip resonators14,15, with potential applications in ultra-fast infor- mation processing, non-reciprocity or microwave to optical transduction. By reducing the number of excitations, MPs find application in the quantum regime e.g., magnon number counting via an electromagnetically coupled superconducting qubit16,17or as a building block for Bell state generation18. The rapidly developing research around polaritons and specifically MPs has so far, been little studied in magnetic vdW materials due to the relatively recent discoveries of long- range magnetic order in vdW systems at the few monolayer regime9,20,21, in addition to its technically challenging real- ization. Stable MP states are formed by strongly coupling the magnetic field oscillation of a resonant photon to the collec- tive magnetization oscillation in a magnetic material. This strong coupling is achieved when the collective coupling rate geffis larger than the average of both system loss rates. In a simplified picture, geffscales linearly with the strength of the oscillating magnetic field of a resonator and the square root number of spins14. For studies involving bulk magnetic mate- rials and low quality and large microwave resonators, strong coupling is achieved when geff=2pis in the MHz range, which is accomplished with relative ease due to the abundance of spins in bulk magnetic materials. A reduction of the bulk di- mensions down from mm to mm and nm scales, the typical lateral dimensions and thickness of vdW material monolay-arXiv:2206.02460v2 [cond-mat.mtrl-sci] 28 Apr 20232 ers, results in a decrease of the coupling strength by at least 6 orders of magnitude. Commonly used microwave resonators are not able to produce strong enough oscillating magnetic fields to compensate for such a reduction in absolute number of spins. Only by advanced resonator design and engineering the regime of strongly coupled MPs in monolayer vdW mag- netic materials can be accomplished, granting access to spin dynamic physics at a true 2d monolayer limit and research on MPs in nano-scale devices where the whole range of on-chip tuning and engineering tools, such as electric fields or device design, are available. Magnons or magnon polaritons have been observed in mag- netic vdW materials, but it had been restricted to either to the optical frequency range22,23or a large thickness limit24,25, re- spectively. Here, we present our attempt of detecting spin dynamics in ultra-thin vdW magnetic materials and the cre- ation of MPs by magnon-photon coupling in the microwave frequency range, using superconducting resonators optimized for increased magnon-photon coupling. By using microwave resonators with a small mode volume, we not only increased its oscillating magnetic field strength but also matched it more efficiently to the size of nanoscale vdW flakes. Our work presents a fundamental cornerstone for a general blueprint for designing and developing magnon-photon hybrids for any type of ultra-thin or monolayer vdW magnetic material, en- abling research on on-chip microwave applications for (quan- tum) information processing. RESULTS In this article, we report on the observation of spin dynam- ics and the creation of MPs at the onset of the high cooper- ativity regime with the vdW ferromagnet CGT of nm scale thickness, demonstrating a pathway towards stable magnon- photon polariton creation. We combine a precise transfer process of exfoliated CGT flakes and high sensitivity su- perconducting resonators, to access and study the dynami- cal response of coupled photon-magnon states in a small- volume (nm-thick and m-sized) CGT flake (illustrated in Fig. 1 (a)). High-quality-factor superconducting lumped el- ement resonators are chosen to be the counterpart due to their extremely small mode volume ( 6000m3) and con- sequently strong oscillating magnetic fields ( B125nT, see SI for resonator quality-factors and B1-field distributions), re- sulting in high spin sensitivities4,26. At cryogenic temper- atures, we perform low-power microwave spectroscopy on multiple resonator-vdW-flake hybrids, covering a frequency range from 12GHz to 18GHz for a variety of thickness. Sam- ples consist of up to 12 resonators on a single chip, all capac- itively coupled to a common microwave transmission line for read-out (see SI for details). Multiple peaks of spin-wave res- onances are observed for each CGT flake measured. The spin- wave modes are closely spaced in frequency and show a large overlap. We employ a semi-optimized fitting model to pro- duce a good estimate for the collective coupling strength and magnetic linewidth. By taking the resonance value of the most prominent peak of each spectrum, we find that all measuredpoints can be fitted very well by a single curve calculated by the Kittel formula with bulk CGT parameters. Furthermore, we extracted the linewidth for the thinnest CGT flake inves- tigated, 11nm or 15 monolayers (ML), the only device ex- hibiting well separated spin-wave modes. This allowed a fully quantitative analysis and we determined an upper limit of the Gilbert damping parameter of 0 :02. This value is comparable to the damping reported for 3d transition metal ferromagnets, suggesting that magnetic vdW flakes have the potential for the fabrication of functional spintronic devices. We investigate the dynamics of nm-thick CGT flakes, us- ing superconducting lumped element resonators made of NbN (see methods for fabrication details and SI and Ref. [28] for more performance details). The advantages of a lumped ele- ment design are the spatial separation of the oscillating mag- netic field B1and electric field E1and the concentration of B1within a narrow wire section of the resonators, as indi- cated in Fig. 1 (a). Additionally, the B1field distribution is homogeneous along the length of the narrow wire section (see finite element simulations in SI). This magnetic-field concen- tration is our primary reason to use this type of resonator in order to reduce the photon mode volume as well as achieve a considerable mode overlap between the resonator photon mode and CGT magnon mode, and consequently, a large cou- pling strength. We therefore transfer CGT flakes onto these 5 μm B0CGTB1a bcE1 0 4 8 x (μm)y (nm) 102030 1240 MCGT CrGeTe B1,extent ≈ 2 μm FIG. 1. Magnon-photon coupling between thin CGT and a super- conducting resonator. a Schematic of a resonator shows the design in detail, indicating the areas of high E1-field (yellow) and B1-field (green) intensities, as well as the orientation of the externally applied field B0. Finally, a schematic zoom in of the section loaded with a CGT flake is shown. The collective coupling between a microwave photon and the magnetization of the CGT is illustrated, as well as the approximate extent of the microwave B1-field. bMicrograph image of a CGT flake transferred onto the narrow section of a resonator. c AFM image of the CGT flake together with a height profile along the blue solid line in the AFM image. The red solid line is a fit to the flake thickness. The results of this resonator are presented in Fig. 2.3 12.8112.8212.83 ω/2π (GHz) 560 580 600 620 Magnetic Field (mT)640 66012.841.0 0.9 0.8|S21|20.7 |S21|21.0 0.9 0.8a b c 580 600 Magnetic Field (mT)620234 κeff/2π (MHz)0510 ωres/2π (MHz) d 640 + 12820 MHz 0.7550 mT 598 mT 614 mT 670 mT 12.81 12.82 12.83 ω/2π (GHz)12.84 FIG. 2. Magnon-photon coupling observed in resonator microwave transmission. a jS21j2as a function static magnetic field B0and frequency, with the microwave transmission encoded in the color. The results are obtained from the resonator shown in Fig 1 (b) and (c), featuring a loaded quality factor of QL=4600. bjS21j2as a function of frequency at fixed magnetic fields, indicated in aby dashed vertical lines. canddResonance frequency wresand effective loss rate keffas a function of magnetic field. Note the multiple resonance peaks, indicating multiple CGT FMRs. The dashed orange lines are results from the semi-optimized fit. dexemplary includes the individual peaks of which the orange dashed lines consists. The green bar in canddhighlights the main mode. narrow sections (Fig. 1 (b)). Details of CGT flake transfers are described in the methods section. Optical imaging and atomic force microscopy (AFM) measurements are used to characterise the size and thickness of the CGT flakes (see Fig. 1 (c)). Measured thicknesses range from 153 23nm down to 111:8nm (15 ML), enabling a thickness dependent study of CGT flakes and their coupling to the resonators. We measured the microwave transmission jS21j2as a func- tion of frequency and externally applied magnetic field B0for each resonator at a temperature of 1 :8K, using a microwave power of approximately 80dBm at the resonator chip. Fig- ure 2 (a) shows the resulting 2D plot of jS21j2for a resonator loaded with a 17nm 0:8nm thick CGT flake (see Fig. 1 (b) and (c) for the respective micrograph and AFM images). A resonator peak can be clearly observed for each magnetic field, with its resonance frequency wresdecreasing with in- creasing magnetic field. The reduction of the frequency is a result of a slow degradation of the superconductivity by B0, which in general exhibits a parabolic dependence29. For 580mTB0630mT the resonator prominence is reduced, highlighted byjS21j2as a function of frequency for four con- stant B0values in Fig. 2 (b). Within this field range, the mode resonance has been modified due to its hybridization with the magnetic modes of the CGT flake. To further quantify the in- teraction, we fit each jS21j2profile by a Fano resonance line- shape (solid orange lines in Fig. 2 (b)) to account for an asym- metric resonance peak due to additional microwave interfer- ence in the circuitry30,31, jS21j2=S0+A(qkeff=2+w wres)2 (keff=2)2+ (w wres)2: (1) Here, S0is the microwave transmission baseline, Athe peak amplitude, qdescribes the asymmetry of the lineshape and keffrepresents the effective loss rate of the hybrid system (seeSI for resonator parameters before and after CGT transfer for all resonators). Figure 2 (c) shows wresof the hybrid system as a function of B0.wresexperiences a dispersive shift when the photon mode and the magnon mode hybridize, indicating an onset of a strong interaction between the two individual systems14,17,32–34. We observe multiple shifts in wres, suggest- ing an interaction of several magnon modes with the resonator in our experiment. Signatures of the resonator–CGT-flake coupling are also characterised by keffof the hybrid system (Fig. 2 (d)). keffis enhanced from the value of the resonator loss rate k0due to an additional loss introduced by the magnon system charac- terized by the loss rate g14,32,35. Consistent with the B0de- pendence of wres,keffshows a rich structure, having its main peak at 598mT, together with less prominent peaks distributed around it. Based on a formalism for coupled-harmonic- oscillator systems in the high cooperativity regime32–34, we use the following to analyse our experimental results with multiple peaks: wres=wres;0+mB2 0++n å k= ng2 eff;kDk D2 k+g2; (2) keff=k0++n å k= ng2 eff;kg D2 k+g2: (3) with the detuning factor for each resonance as Dk= gCGTmB ¯h B0 BFMR ;k . Here, wres;0is the resonator resonance frequency at B0=0T and mrepresents the curvature of the resonance frequency decrease due to the applied magnetic field. BFMR ;kis the CGT FMR field, gCGT the g-factor of CGT and geff;kgives the collective coupling strength between photon and magnon mode. The summation is over all reso- nance modes kpresent on the low or high field (frequency)4 side of the main resonance mode, where ngives the number of modes on one side. For simplicity, we assume a symmet- ric distribution of modes about the main mode. The large number of multiple modes and their strong overlap prevent a reliable application of a fully optimized fit to the data, due to the large number of free parameters required. In an ef- fort to gain a good estimate of the model parameters we ap- ply the model functions Eq. (2) and (3) in a two-step semi- optimized fashion (see SI for details). With this approach, we arrive at a model in good agreement with wresandkeff(see orange dashed lines in Fig. 2 (c), (d), exemplary showing the individual peaks of the orange dashed line in Fig. 2 (d) and the SI for additional results and data). We can reproduce the data using g=2p=94:035:95MHz and a collective cou- pling strength of the main mode of 13 :251MHz. Together with k0=2p=1:40:02MHz the system resides at the onset of the high cooperativity regime, classified by the cooperativ- ityC=g2 eff=k0g=1:3>113,32. In this regime, magnon polari- tons are created and coherently exchange excitations between magnons and resonator photons on a rate given by geff. The created MPs are, however, short lived and the excitations pre- 100 200 500 700 Resonance Field BFMR (mT)300 400 600 0051015ωFMR/2π (GHz)500 Resonance Field BFMR (mT)600 7001518ωFMR/2π (GHz) 12a b 11 31 51 71 91 111 131151Flake Thickness (nm) FIG. 3. Summary of CGT-FMR conditions. a Extracted CGT res- onance fields and frequencies from the set of resonators loaded with CGT flakes of different thickness. Resonance values are taken from the most prominent peaks in keff. The solid curve is calculated us- ing the Kittel formalism presented in10, using same parameters, with gCGT =2:18,m0Ms=211:4mT and Ku=3:84104J=m3.bWider magnetic field range of awhere the CGT flake thickness for the dif- ferent symbols is indicated by the color gradient given in a.dominately dissipate in the magnonic system, as geffg. Our analysis suggests that the separation of the different FMR modes is of the same order of magnitude as the loss rate (see SI for additional data). We consider that these are from standing spin wave resonances, commonly observed for thin magnetic films12and with one reported observation in bulk of the vdW material CrI 338. In thin-film magnets under a static magnetic field applied in-plane, the magnetic-dipole interac- tion generates two prominent spin wave branches for an in- plane momentum, the backward volume spin wave (BVMSW) and magnetostatic surface spin wave (MSSW) modes39,40. These spin wave modes have different dispersion relations, having higher (MSSWs) and lower (BVMSWs) resonance frequencies with respect to that of the uniform FMR mode. We calculate the distance of these standing spin-wave modes based on magnetic parameters of bulk CGT as well as the lat- eral dimensions of the flakes (see SI for more details). We can find spin waves having a frequency separation within 100MHz and 200MHz (3 :3mT to 6 :6mT in magnetic field units), which are consistent with our experimental observa- tion in terms of its mode separation. However, the irregular shape of the CGT flakes renders exact calculations of spin wave mode frequencies very challenging. We also consid- ered a possibility that each layer of CGT might have different magnetic parameters (e.g. chemical inhomogeneity), and thus producing different individual resonance modes. Our numer- ical simulations based on atomistic spin dynamics14,15rule out this possibility, as resonance modes from individual lay- ers average to a single mode as soon as a fraction of 10% of inter-layer exchange coupling is introduced (see SI for more details). Therefore, we speculate that the multiple mode na- ture we observe in our experiments is likely originating from intrinsic properties of the CGT flakes. Figure 3 shows the extracted wFRM as a function of BFMR for each resonator–CGT-flake hybrid. The experimental val- ues are in excellent agreement with a curve calculated by the Kittel equation with magnetic parameters for bulk CGT10, from which the data exhibits a standard deviation of less than 5%. This agreement, achieved by independent characteri- zations of 23 CGT flakes measured by superconducting res- onators, is experimental evidence that the magnetic parame- ters that determine the dispersion of wFRM (BFMR), i.e. the CGT g-factor gCGT, saturation magnetization Msand uniaxial anisotropy Ku, exhibit little thickness dependence in exfoli- ated CGT flakes, and are not disturbed by the transfer onto the resonator structure. We note, that this demonstrates that vdW magnetic materials are particularly attractive for device applications, as they are less prone for contamination from exfoliation. Finally, we present our analysis of kefffor a resonator with a 111:8nm CGT flake in Fig.4. With the thickness of a single layer of CGT being 0 :7nm9, this flake consists of 15 monolayers and is the thinnest in our series. Figure 4 (a) and (b) show wresandkeffas a function of B0, respectively. While the response of the CGT flake shows a prominent signature inkeff, the CGT FMR is considerably more subtle in wres. This highlights the excellent sensitivity of the high-Q super- conducting resonators in our study. kefffeatures five well-5 separated peaks with the main peak at B0=547mT, which enables us to perform a single-peak fully optimized analy- sis for each, in contrast to our multi-step analysis for the re- mainder of the devices. We assume the additional peaks are BVMSW modes, as discussed in the previous section. How- ever, the splitting is about four times larger than compared to all other investigated devices, which would result in a signifi- cantly shorter wavelength. Thickness steps can lead to a wave- length down-conversion13, however, due to the irregular shape andB1inhomogeneities it is difficult to exactly calculate the spin wave frequencies (see SI for further details). From the main peak profile, we extract geff=2p=3:610:09MHz, g=2p=126:268:5MHz and k0=2p=0:920:05MHz. We compare the experimental value of geffwith a numerically cal- culated geff;simu, using the dimensions of the CGT flake de- termined by AFM measurements (see SI for details). The calculation yields geff;simu=2p=8:94MHz, lying within the same order of magnitude. The overestimation is likely due to in-perfect experimental conditions, like non-optimal place- ment of the flake, uncertainties in the thickness and dimen- sion determination as well as excluding the additional modes in the calculation (see SI). With ggeffandC=0:11, the hybrid system is in the weak coupling regime13, but due to the highly sensitive resonator with its small k0the response from the magnon system can still be detected. With the ex- tracted g=2pwe can give an upper limit of the Gilbert damp- ing in CGT, by calculating aupper =g=wFMR. We find aupper as 0:0210:002, which is comparable to other transition metal magnetic materials44, and is in very good agreement with a previously reported effective Gilbert damping parameter de- termined by laser induced magnetization dynamics45. Here, we emphasise that the actual Gilbert damping value is lower due to a finite, extrinsic inhomogeneous broadening contribu- tion. We further use these results to benchmark the sensitivity of our measurement techniques. The detection limit is given by comparing the main peak height characterised by g2 eff=g and the median noise amplitude which is 18kHz in Fig. 4 (b) where g2 eff=2pg= 103 kHz. By assuming the same lateral di- mensions and scale the thickness down to a single monolayer, while keeping gconstant, we calculate the expected signal re- duction numerically by geff;simu;1ML=geff;simu;15ML to 0.26. We obtain (0:26geff)2=2pg=7kHz for the monolayer limit. Al- though this suggests the noise amplitude is greater than the expected peak amplitude, we can overturn this condition by improving the coupling strength by optimising the resonator design, enhancing the exfoliation and flake transfer as well as by reducing the noise level by averaging a number of mul- tiple scans. Superconducting resonators with mode volumes of about 10 m3have been realised46, a reduction of 2 orders of magnitude compared to our current design. This would translate to an order of magnitude improvement in geff. Fur- thermore, this flake covers about 4% of the resonator. By assuming maximised coverage a 5 times enhancement of geff can be achieved. Both approaches would make the detection of monolayer flakes possible. In summary, we provide the first demonstration of photon- magnon coupling between a superconducting resonator and 520 540 600 6400.900.951.001.05 κeff/2π (MHz) Magnetic Field (mT)560 580 620 ωres/2π (MHz) 122801229012300a bFIG. 4. Magnon-photon coupling for the thinnest CGT flake. a Resonance frequency wresandbeffective loss rate keffas a func- tion of magnetic field of a resonator loaded with the thinnest CGT, consisting of 15 ML. The resonator’s loaded quality factor is 6938. The solid orange lines are results a fit to Eq.(2) and (3), respectively. The errorbars in brepresent the standard deviation from the Fano resonance lineshape fit to the resonator transmission. nm-thick vdW flakes of CGT, using a total of 23 devices with different CGT flakes of thickness from 153nm down to 11nm. By employing a coupled-harmonic-oscillator model, we extract the coupling strength, magnetic resonance field and relaxation rates for both photon and magnon modes in our devices. From our semi-broadband experiments, we find that the magnetic properties of exfoliated CGT flakes are ro- bust against the transfer process, with a standard deviation of less than 5% to expected resonance values from bulk param- eters. Notably, this suggests that vdW magnetic materials can be pre-screened at bulk to identify the most promising mate- rial for few layer device fabrication. The upper limit of the Gilbert damping in the 15 ML thick CGT flake is determined to be 0 :021, which is comparable to commonly used ferro- magnetic thin-films such as NiFe and CoFeB and thus mak- ing CGT attractive for similar device applications. We high- light that the damping parameter is key in precessional mag- netisation switching47,48, auto-oscillations by dc currents49,50, and comprehensive spin-orbit transport in vdW magnetic sys- tems51. The presented techniques are readily transferable to other vdW magnetic systems to study spin dynamics in atomically-thin crystalline materials. While creating stable magnon polaritons is still an open challenge due to the large loss rate gof the CGT magnon system, this work offers an important approach towards its achievement. There are still potential improvements to the measurement sensitivity such as resonator mode volume reduction by introducing nm scale constrictions52,53and use of exfoliation/transfer techniques to produce larger flakes to enhance the mode overlap (hence cou-6 pling strength)54,55. With concerted efforts, the formation of magnon polaritons in few layers vdW materials will become feasible. METHODS Superconducting Resonators: The resonators were fab- ricated by direct laser writing and a metal lift-off process. The individual 5mm 5mm chips are scribed from an in- trinsic, high resistivity ( r>5000Wcm) n-type silicon wafer of 250 m thickness. For a well defined lift-off, we use a double photoresist layer of LOR and SR1805. The resonator structures are transferred into the resist by a Heidelberg Di- rect Writer system. After development, 50nm NbN are de- posited by magnetron sputtering in a SVS6000 chamber, at a base pressure of 7 10 7mbar, using a sputter power of 200W in an 50:50 Ar/N atmosphere held at 5 10 3mbar, with both gas flows set to 50 SCCM28. Finally, the lift-off is done in a 1165 solvent to release the resonator structures. CGT Crystal Growth: CGT crystals used in this study were grown via chemical vapour transport. To this end, high- purity elemental precursors of Cr (chips, 99:995%), Ge (powder,99:999%), and Te (shots, 99 :999%) were mixed in the molar weight ratio Cr:Ge:Te = 10:13.5:76.5, loaded into a thick-wall quartz ampule and sealed under the vacuum of 10 5mbar. Then, the ampule was loaded into a two-zone furnace, heated up and kept at 950C for 1 week to homog- enize the precursors. To ensure high-quality growth, the am- pule was slowly cooled (0 :4C=h) maintaining a small tem- perature gradient between the opposite ends of the ampule. Once the ampule reached 500C, the furnace was turned off allowing the ampule to cool down to room temperature nat- urally. The large ( 1cm) single-crystalline flakes were ex- tracted from the excess tellurium and stored in the inert envi- ronment. CGT Flake Transfer: Devices for this study were made via transfer of single-crystalline thin flakes on top of the super- conducting resonators. The flakes were first exfoliated from bulk crystals on the clean surface of a home-cured PDMS (polydimethylsiloxane, Sylgard 184) substrate. The thickness of the CGT flakes on PDMS was estimated through the con- trast variation with transmission optical microscopy. Then, the selected flake was transferred to a resonator. The trans- fer was performed in air at room temperature. To minimize the air exposure, the entire process of exfoliation, inspection and transfer was reduced to 10-15 min per resonator. For the flakes thicker than 50nm, the strong optical absorption of CGT prevented the accurate thickness estimation with optical contrast. For those flakes, the thickness was estimated via a quick AFM scan performed on the PDMS substrate before the transfer step. Ready devices were stored in inert conditions. DATA AVAILABILITY The data that support the findings of this study are available within the paper, Supplemental Material and from the corre-sponding authors upon reasonable request. REFERENCES 1A. K. Geim and I. V . Grigorieva, “Van der waals heterostructures,” Nature , vol. 499, pp. 419–425, Jul 2013. 2K. S. Novoselov, A. Mishchenko, A. Carvalho, and A. H. C. Neto, “2d materials and van der waals heterostructures,” Science , vol. 353, no. 6298, p. 9439, 2016. 3Q. H. Wang, A. Bedoya-Pinto, M. Blei, A. H. Dismukes, A. Hamo, S. Jenk- ins, M. Koperski, Y . Liu, Q.-C. Sun, E. J. Telford, H. H. Kim, M. Augustin, U. V ool, J.-X. Yin, L. H. Li, A. Falin, C. R. Dean, F. Casanova, R. F. L. Evans, M. Chshiev, A. Mishchenko, C. Petrovic, R. He, L. Zhao, A. W. Tsen, B. D. Gerardot, M. Brotons-Gisbert, Z. Guguchia, X. Roy, S. Ton- gay, Z. Wang, M. Z. Hasan, J. Wrachtrup, A. Yacoby, A. Fert, S. Parkin, K. S. Novoselov, P. Dai, L. Balicas, and E. J. G. Santos, “The magnetic genome of two-dimensional van der waals materials,” ACS Nano , 04 2022. 4C. Tang, L. Alahmed, M. Mahdi, Y . Xiong, J. Inman, N. J. McLaugh- lin, C. Zollitsch, T. H. Kim, C. R. Du, H. Kurebayashi, E. J. G. Santos, W. Zhang, P. Li, and W. Jin, “Ferromagnetic resonance in two-dimensional van der waals magnets: A probe for spin dynamics,” 2023. 5D. N. Basov, M. M. Fogler, and F. J. G. de Abajo, “Polaritons in van der waals materials,” Science , vol. 354, no. 6309, p. 1992, 2016. 6T. Low, A. Chaves, J. D. Caldwell, A. Kumar, N. X. Fang, P. Avouris, T. F. Heinz, F. Guinea, L. Martin-Moreno, and F. Koppens, “Polaritons in layered two-dimensional materials,” Nature Materials , vol. 16, pp. 182– 194, Feb 2017. 7Z. Fei, A. S. Rodin, G. O. Andreev, W. Bao, A. S. McLeod, M. Wagner, L. M. Zhang, Z. Zhao, M. Thiemens, G. Dominguez, M. M. Fogler, A. H. C. Neto, C. N. Lau, F. Keilmann, and D. N. Basov, “Gate-tuning of graphene plasmons revealed by infrared nano-imaging,” Nature , vol. 487, pp. 82–85, Jul 2012. 8J. Chen, M. Badioli, P. Alonso-González, S. Thongrattanasiri, F. Huth, J. Osmond, M. Spasenovi ´c, A. Centeno, A. Pesquera, P. Godignon, A. Zu- rutuza Elorza, N. Camara, F. J. G. de Abajo, R. Hillenbrand, and F. H. L. Koppens, “Optical nano-imaging of gate-tunable graphene plasmons,” Na- ture, vol. 487, pp. 77–81, Jul 2012. 9X. Liu, T. Galfsky, Z. Sun, F. Xia, E.-c. Lin, Y .-H. Lee, S. Kéna-Cohen, and V . M. Menon, “Strong light–matter coupling in two-dimensional atomic crystals,” Nature Photonics , vol. 9, pp. 30–34, Jan 2015. 10I. A. Verzhbitskiy, H. Kurebayashi, H. Cheng, J. Zhou, S. Khan, Y . P. Feng, and G. Eda, “Controlling the magnetic anisotropy in Cr 2Ge2Te6by electro- static gating,” Nature Electronics , vol. 3, no. 8, pp. 460–465, 2020. 11B. Zare Rameshti, S. Viola Kusminskiy, J. A. Haigh, K. Usami, D. Lachance-Quirion, Y . Nakamura, C.-M. Hu, H. X. Tang, G. E. W. Bauer, and Y . M. Blanter, “Cavity magnonics,” Physics Reports , vol. 979, pp. 1– 61, 2022. 12D. D. Awschalom, C. R. Du, R. He, F. J. Heremans, A. Hoffmann, J. Hou, H. Kurebayashi, Y . Li, L. Liu, V . Novosad, J. Sklenar, S. E. Sullivan, D. Sun, H. Tang, V . Tyberkevych, C. Trevillian, A. W. Tsen, L. R. Weiss, W. Zhang, X. Zhang, L. Zhao, and C. W. Zollitsch, “Quantum engineer- ing with hybrid magnonic systems and materials,” IEEE Transactions on Quantum Engineering , vol. 2, pp. 1–36, 2021. 13X. Zhang, C.-L. Zou, L. Jiang, and H. X. Tang, “Strongly coupled magnons and cavity microwave photons,” Physical Review Letters , vol. 113, p. 156401, Oct 2014. 14H. Huebl, C. W. Zollitsch, J. Lotze, F. Hocke, M. Greifenstein, A. Marx, R. Gross, and S. T. B. Goennenwein, “High cooperativity in coupled mi- crowave resonator ferrimagnetic insulator hybrids,” Physical Review Let- ters, vol. 111, p. 127003, Sep 2013. 15Y .-P. Wang, J. W. Rao, Y . Yang, P.-C. Xu, Y . S. Gui, B. M. Yao, J. Q. You, and C.-M. Hu, “Nonreciprocity and unidirectional invisibility in cavity magnonics,” Physical Review Letters , vol. 123, p. 127202, Sep 2019. 16Y . Tabuchi, S. Ishino, A. Noguchi, T. Ishikawa, R. Yamazaki, K. Usami, and Y . Nakamura, “Coherent coupling between a ferromagnetic magnon and a superconducting qubit,” Science , vol. 349, no. 6246, pp. 405–408, 2015.7 17D. Lachance-Quirion, Y . Tabuchi, A. Gloppe, K. Usami, and Y . Nakamura, “Hybrid quantum systems based on magnonics,” Applied Physics Express , vol. 12, p. 070101, jun 2019. 18H. Y . Yuan, Y . Cao, A. Kamra, R. A. Duine, and P. Yan, “Quantum magnonics: When magnon spintronics meets quantum information sci- ence,” Physics Reports , vol. 965, pp. 1–74, 2022. 19C. Gong, L. Li, Z. Li, H. Ji, A. Stern, Y . Xia, T. Cao, W. Bao, C. Wang, Y . Wang, Z. Q. Qiu, R. J. Cava, S. G. Louie, J. Xia, and X. Zhang, “Discov- ery of intrinsic ferromagnetism in two-dimensional van der waals crystals,” Nature , vol. 546, pp. 265–269, Jun 2017. 20B. Huang, G. Clark, E. Navarro-Moratalla, D. R. Klein, R. Cheng, K. L. Seyler, D. Zhong, E. Schmidgall, M. A. McGuire, D. H. Cobden, W. Yao, D. Xiao, P. Jarillo-Herrero, and X. Xu, “Layer-dependent ferromagnetism in a van der waals crystal down to the monolayer limit,” Nature , vol. 546, pp. 270–273, Jun 2017. 21J.-U. Lee, S. Lee, J. H. Ryoo, S. Kang, T. Y . Kim, P. Kim, C.-H. Park, J.- G. Park, and H. Cheong, “Ising-type magnetic ordering in atomically thin FePS 3,”Nano Letters , vol. 16, pp. 7433–7438, Dec. 2016. 22X.-X. Zhang, L. Li, D. Weber, J. Goldberger, K. F. Mak, and J. Shan, “Gate- tunable spin waves in antiferromagnetic atomic bilayers,” Nature Materials , vol. 19, no. 8, pp. 838–842, 2020. 23J. Cenker, B. Huang, N. Suri, P. Thijssen, A. Miller, T. Song, T. Taniguchi, K. Watanabe, M. A. McGuire, D. Xiao, and X. Xu, “Direct observation of two-dimensional magnons in atomically thin CrI3,” Nature Physics , vol. 17, no. 1, pp. 20–25, 2021. 24S. Mandal, L. N. Kapoor, S. Ghosh, J. Jesudasan, S. Manni, A. Thamizhavel, P. Raychaudhuri, V . Singh, and M. M. Deshmukh, “Copla- nar cavity for strong coupling between photons and magnons in van der waals antiferromagnet,” Applied Physics Letters , vol. 117, p. 263101, Dec. 2020. 25Q. Zhang, Y . Sun, Z. Lu, J. Guo, J. Xue, Y . Chen, Y . Tian, S. Yan, and L. Bai, “Zero-field magnon-photon coupling in antiferromagnet CrCl3,” Applied Physics Letters , vol. 119, p. 102402, Sept. 2021. 26C. Eichler, A. J. Sigillito, S. A. Lyon, and J. R. Petta, “Electron spin reso- nance at the level of 104spins using low impedance superconducting res- onators,” Physical Review Letters , vol. 118, p. 037701, Jan 2017. 27S. Weichselbaumer, P. Natzkin, C. W. Zollitsch, M. Weiler, R. Gross, and H. Huebl, “Quantitative modeling of superconducting planar resonators for electron spin resonance,” Physical Review Applied , vol. 12, p. 024021, Aug 2019. 28C. W. Zollitsch, J. O’Sullivan, O. Kennedy, G. Dold, and J. J. L. Morton, “Tuning high-Q superconducting resonators by magnetic field reorienta- tion,” AIP Advances , vol. 9, p. 125225, Dec. 2019. 29J. E. Healey, T. Lindstroem, M. S. Colclough, C. M. Muirhead, and A. Y . Tzalenchuk, “Magnetic field tuning of coplanar waveguide resonators,” Ap- plied Physics Letters , vol. 93, p. 043513, July 2008. 30U. Fano, “Effects of configuration interaction on intensities and phase shifts,” Physical Review , vol. 124, pp. 1866–1878, Dec 1961. 31M. S. Khalil, M. J. A. Stoutimore, F. C. Wellstood, and K. D. Osborn, “An analysis method for asymmetric resonator transmission applied to super- conducting devices,” Journal of Applied Physics , vol. 111, no. 5, 2012. 32P. F. Herskind, A. Dantan, J. P. Marler, M. Albert, and M. Drewsen, “Real- ization of collective strong coupling with ion coulomb crystals in an optical cavity,” Nature Physics , vol. 5, no. 7, pp. 494–498, 2009. 33P. Bushev, A. K. Feofanov, H. Rotzinger, I. Protopopov, J. H. Cole, C. M. Wilson, G. Fischer, A. Lukashenko, and A. V . Ustinov, “Ultralow-power spectroscopy of a rare-earth spin ensemble using a superconducting res- onator,” Physical Review B , vol. 84, p. 060501, Aug 2011. 34S. Khan, O. Lee, T. Dion, C. W. Zollitsch, S. Seki, Y . Tokura, J. D. Breeze, and H. Kurebayashi, “Coupling microwave photons to topological spin tex- tures in Cu 2OSeO 3,”Physical Review B , vol. 104, p. L100402, Sep 2021. 35Y . Tabuchi, S. Ishino, T. Ishikawa, R. Yamazaki, K. Usami, and Y . Naka- mura, “Hybridizing ferromagnetic magnons and microwave photons in the quantum limit,” Physical Review Letters , vol. 113, p. 083603, Aug 2014. 36S. Khan, C. W. Zollitsch, D. M. Arroo, H. Cheng, I. Verzhbitskiy, A. Sud, Y . P. Feng, G. Eda, and H. Kurebayashi, “Spin dynamics study in lay- ered van der waals single-crystal Cr 2Ge2Te6,”Physical Review B , vol. 100, p. 134437, Oct 2019. 37A. A. Serga, A. V . Chumak, and B. Hillebrands, “Yig magnonics,” Journal of Physics D: Applied Physics , vol. 43, no. 26, p. 264002, 2010.38L. N. Kapoor, S. Mandal, P. C. Adak, M. Patankar, S. Manni, A. Thamizhavel, and M. M. Deshmukh, “Observation of standing spin waves in a van der waals magnetic material,” Advanded Materials , vol. 33, p. 2005105, Jan. 2021. 39B. A. Kalinikos and A. N. Slavin, “Theory of dipole-exchange spin wave spectrum for ferromagnetic films with mixed exchange boundary condi- tions,” Journal of Physics C: Solid State Physics , vol. 19, pp. 7013–7033, dec 1986. 40U. K. Bhaskar, G. Talmelli, F. Ciubotaru, C. Adelmann, and T. Devolder, “Backward volume vs damon-eshbach: A traveling spin wave spectroscopy comparison,” Journal of Applied Physics , vol. 127, p. 033902, Jan. 2020. 41D. A. Wahab, M. Augustin, S. M. Valero, W. Kuang, S. Jenkins, E. Coron- ado, I. V . Grigorieva, I. J. Vera-Marun, E. Navarro-Moratalla, R. F. Evans, et al. , “Quantum rescaling, domain metastability, and hybrid domain-walls in 2d CrI3 magnets,” Advanced Materials , vol. 33, no. 5, p. 2004138, 2021. 42A. Kartsev, M. Augustin, R. F. Evans, K. S. Novoselov, and E. J. G. San- tos, “Biquadratic exchange interactions in two-dimensional magnets,” npj Computational Materials , vol. 6, no. 1, pp. 1–11, 2020. 43J. Stigloher, T. Taniguchi, M. Madami, M. Decker, H. S. Körner, T. Moriyama, G. Gubbiotti, T. Ono, and C. H. Back, “Spin-wave wave- length down-conversion at thickness steps,” Applied Physics Express , vol. 11, p. 053002, apr 2018. 44S. Mankovsky, D. Ködderitzsch, G. Woltersdorf, and H. Ebert, “First- principles calculation of the gilbert damping parameter via the linear re- sponse formalism with application to magnetic transition metals and al- loys,” Physical Review B , vol. 87, p. 014430, Jan 2013. 45T. Zhang, Y . Chen, Y . Li, Z. Guo, Z. Wang, Z. Han, W. He, and J. Zhang, “Laser-induced magnetization dynamics in a van der waals ferromagnetic Cr2Ge2Te6nanoflake,” Applied Physics Letters , vol. 116, p. 223103, June 2020. 46S. Probst, A. Bienfait, P. Campagne-Ibarcq, J. J. Pla, B. Albanese, J. F. Da Silva Barbosa, T. Schenkel, D. Vion, D. Esteve, K. Mølmer, J. J. L. Morton, R. Heeres, and P. Bertet, “Inductive-detection electron-spin reso- nance spectroscopy with 65 spins/Hz sensitivity,” Applied Physics Letters , vol. 111, p. 202604, Nov. 2017. 47G. E. Rowlands, C. A. Ryan, L. Ye, L. Rehm, D. Pinna, A. D. Kent, and T. A. Ohki, “A cryogenic spin-torque memory element with precessional magnetization dynamics,” Scientific Reports , vol. 9, no. 1, p. 803, 2019. 48A. Meo, J. Chureemart, R. W. Chantrell, and P. Chureemart, “Magnetisation switching dynamics induced by combination of spin transfer torque and spin orbit torque,” Scientific Reports , vol. 12, no. 1, p. 3380, 2022. 49K. Wagner, A. Smith, T. Hache, J.-R. Chen, L. Yang, E. Montoya, K. Schultheiss, J. Lindner, J. Fassbender, I. Krivorotov, and H. Schultheiss, “Injection locking of multiple auto-oscillation modes in a tapered nanowire spin hall oscillator,” Scientific Reports , vol. 8, no. 1, p. 16040, 2018. 50M. Haidar, A. A. Awad, M. Dvornik, R. Khymyn, A. Houshang, and J. Åk- erman, “A single layer spin-orbit torque nano-oscillator,” Nature Commu- nications , vol. 10, no. 1, p. 2362, 2019. 51H. Kurebayashi, J. H. Garcia, S. Khan, J. Sinova, and S. Roche, “Mag- netism, symmetry and spin transport in van der waals layered systems,” Nature Reviews Physics , vol. 4, pp. 150–166, Mar 2022. 52L. McKenzie-Sell, J. Xie, C.-M. Lee, J. W. A. Robinson, C. Ciccarelli, and J. A. Haigh, “Low-impedance superconducting microwave resonators for strong coupling to small magnetic mode volumes,” Physical Review B , vol. 99, p. 140414, Apr 2019. 53I. Gimeno, W. Kersten, M. C. Pallarés, P. Hermosilla, M. J. Martínez- Pérez, M. D. Jenkins, A. Angerer, C. Sánchez-Azqueta, D. Zueco, J. Majer, A. Lostao, and F. Luis, “Enhanced molecular spin-photon coupling at su- perconducting nanoconstrictions,” ACS Nano , vol. 14, pp. 8707–8715, July 2020. 54Y . Huang, Y .-H. Pan, R. Yang, L.-H. Bao, L. Meng, H.-L. Luo, Y .-Q. Cai, G.-D. Liu, W.-J. Zhao, Z. Zhou, L.-M. Wu, Z.-L. Zhu, M. Huang, L.-W. Liu, L. Liu, P. Cheng, K.-H. Wu, S.-B. Tian, C.-Z. Gu, Y .-G. Shi, Y .-F. Guo, Z. G. Cheng, J.-P. Hu, L. Zhao, G.-H. Yang, E. Sutter, P. Sutter, Y .-L. Wang, W. Ji, X.-J. Zhou, and H.-J. Gao, “Universal mechanical exfoliation of large-area 2d crystals,” Nature Communications , vol. 11, no. 1, p. 2453, 2020. 55J. Zhou, C. Zhang, L. Shi, X. Chen, T. S. Kim, M. Gyeon, J. Chen, J. Wang, L. Yu, X. Wang, K. Kang, E. Orgiu, P. Samorì, K. Watanabe, T. Taniguchi, K. Tsukagoshi, P. Wang, Y . Shi, and S. Li, “Non-invasive digital etching1 of van der waals semiconductors,” Nature Communications , vol. 13, no. 1, p. 1844, 2022. ACKNOWLEDGMENT This study is supported by EPSRC on EP/T006749/1 and EP/V035630/1. G.E. acknowledges support from the Min- istry of Education (MOE), Singapore, under AcRF Tier 3 (MOE2018-T3-1-005) and the Singapore National Research Foundation for funding the research under medium-sized cen- tre program. E.J.G.S. acknowledges computational resources through CIRRUS Tier-2 HPC Service (ec131 Cirrus Project) at EPCC (http://www.cirrus.ac.uk) funded by the University of Edinburgh and EPSRC (EP/P020267/1); ARCHER UK Na- tional Supercomputing Service (http://www.archer.ac.uk) via Project d429. E.J.G.S acknowledges the Spanish Ministry of Science’s grant program “Europa-Excelencia” under grant number EUR2020-112238, the EPSRC Early Career Fellow- ship (EP/T021578/1), and the University of Edinburgh for funding support. D.S. acknowledges EPSRC funding through the Centre for Doctoral Training in Advanced Characteri- sation of Materials (EP/L015277/1) and European Union’s Horizon 2020 Research and Innovation program under grantagreement GrapheneCore3, number 881603 and the Depart- ment for Business, Energy and Industrial Strategy through the NPL Quantum Program. AUTHOR CONTRIBUTION C.W.Z, S.K. and H.K. conceived the experimental project. Resonator design and optimization was done by J.O’S., O.W.K, C.W.Z and supervised by J.J.L.M. Resonator fabri- cation and characterization was done by C.W.Z. CGT crystals were grown by I.A.V . and exfoliated and transferred by I.A.V . and N.V .T.T. and supervised by E.G.. D.S. measured AFM on the CGT flakes on the resonators. C.W.Z. performed the ex- periments and the data analysis with input from S.K. and H.K. Atomistic spin dynamics simulations were carried out by M.S. supervised by E.J.G.S.. C.W.Z., M.S., I.A.V . and H.K. wrote the manuscript with input from all authors. COMPETING INTERESTS The Authors declare no conflict of interests. Supplemental Material - Probing spin dynamics of ultra-thin van der Waals magnets via photon-magnon coupling I. MICROWAVE SETUP AND MEASUREMENT VNA MW out MW in -20 dB+32 dB DUT B0Cryostat FIG. S1. Microwave delivery and detection setup. Schematic of the microwave delivery and detection circuit. The image shows the coplanar waveguide transmission line. A resonator chip is placed on top of the transmission line for read out. On the right, a schematic layout of the resonators on a single chip is shown.2 Figure S1 shows a schematic of the used microwave measurement setup. We are using a Keysight E5071C vector network analyzer (VNA) to deliver and detect microwaves. The VNA is connected to a low temperature probe, fitted into a closed cycle helium cryostat and cooled to a base temperature of about 1 :8K. The microwave signal is transmitted into the cryostat and is attenuated by 20dB. The attenuator is positioned just before the sample box and provides a thermal anchoring for the center conductor of the coaxial cable to minimize the thermal load onto the sample. The output line is equipped with a Low Noise Factory LNC6_20C cryogenic amplifier, operating between 6 20GHz with an average amplification of +32dB. The transmitted and amplified signal is finally detected by the VNA. Figure S1 also shows an image of the coplanar waveguide transmission line PCB, loaded with a resonator ship, of which a schematic shows the resonator layout on a single chip. The resonators on the chip are capacitively coupled to the transmission line PCB. Upon resonance the transmission through the PCB is reduced, indicating the resonator resonance. The cryostat is equipped with a mechanical rotation stage and prior to the measurements the superconducting resonators are carefully aligned to the externally applied static magnetic field B0, such that the field is in the plane of the superconductor and along the narrow section of the resonators. Figure S2 shows the raw uncalibrated microwave transmission, ranging from 10GHz to 18GHz. The transmission is domi- nated by imperfections in our microwave circuitry, masking the small signals from the superconducting resonators. Thus, we performed a simple thru calibration of the microwave transmission to remove contributions from the setup, prior each magnetic field dependent measurement. Here, we exploit the magnetic field tunability of our superconducting resonators. Before calibra- tion, we set the frequency range of the measurement. We change the applied magnetic field such that the resonator’s resonance frequency is tuned out of the set frequency range. With a frequency window just showing the transmission of the setup we perform the thru calibration. After calibration we set the magnetic field back to its starting value, resulting in a background corrected spectrum with just the resonator feature on it. 10 11 12 13 14 15 16 17 18-1010 0 Raw Transmission |S21| (dB) Frequency (GHz) FIG. S2. Raw broadband microwave transmission signal. Logarithmic microwave transmission jS21jas a function of frequency between 10GHz and 18GHz at a temperature of 1 :8K. II. RESONATOR CHARACTERIZATION In this study, we fabricate twelve superconducting lumped element resonators on each of three resonator chips were fabricated using the same design (see schematic Fig. 1 (a) in the main text). Prior to transfer of the CGT flakes, we characterized the res- onators at a temperature of 1 :8K and zero applied magnetic field, using microwave powers of about 80dBm at the resonators, which is well below the bifurcation limit starting above 60dBm. Due to finite fabrication tolerances the resonator parameters have some variation, while some didn’t work at all. However, the targeted resonance frequencies are well reproducible and very similar for the 3 different chips. We compare the resonator parameters before and after transfer of the CGT flakes and collate the parameters in Tab. I. Note, the resonator parameters with the CGT flakes on were obtained with a static magnetic field applied in the plane of the superconductor, but far detuned from the CGT FMR. In addition, we add the respective thickness of the flake on each resonator, acquired from AFM measurements. Here, we give the values of the thickest region of a given flake on a resonator, as the thickest region will dominate the FMR signal. Due to the arbitrary shape of exfoliated flakes, some exhibit regions of different thickness, as seen e.g. in Fig. S5 (h) and (i).3 TABLE I. Resonator Parameters Chip Number wres;before (MHz) QL;before wres;after(MHz) QL;after CGT Thickness (nm) 1 12165 1978 12063 5733 16.21.3 1 13303 7357 13177 4950 - 1 13968 5575 13860 4679 49.43.5 1 14184 6492 14048 5627 153.123.3 1 16648 6606 16470 5021 23.52.5 1 17431 3215 17237 6826 23.86.4 1 17959 7595 17790 3963 26.24.1 2 12285 360 12153 7135 49.19.1 2 12669 3600 12548 6693 102.85.6 2 12782 3448 12648 6557 105.93.9 2 13393 4643 13244 4501 34.44.1 2 13760 6858 13620 5488 95.95.9 2 14395 9048 14201 4139 36.74.3 2 16075 7283 - - - 2 17048 6541 16811 4241 75.55.4 3 12043 6114 11899 6044 59.732.8 3 12456 2716 12314 6938 11.41.8 3 12996 5828 12848 4600 170.8 3 13422 6517 13272 5461 89.87.5 3 13719 6800 13582 6608 - 3 14238 9184 14064 5420 73.58.4 3 15390 8680 15219 6030 30.54.2 3 15821 2386 15604 4769 33.19.9 3 16430 7518 16193 5780 30.138.1 3 17308 6521 17054 5569 137.93.4 3 18111 3542 17870 4643 50.26.9 III. RESONATOR AND COUPLING SIMULATION We use finite element and numerical simulations to optimize our resonator design. Key requirements of our resonators are a strong resilience to externally applied static magnetic fields and a small mode volume. To achieve a large field resilience we reduced the area of the resonator to minimize effects of the magnetic field on the superconducting film. Further, we designed the resonators such that they act as lumped element resonators. Here, the resonance frequency is given by the total capacitance and inductance of the structure, with wres=1=p LC, analogues to a parallel LC circuit. This allows us to locally separate oscillating electric and magnetic fields and also to concentrate the magnetic fields in more confined regions, resulting in very small mode volumes. To verify the lumped element nature of our resonators we performed finite element simulations, using CST Microwave Studio. Figure S3 shows the resulting magnitude of the E-field (left side) and H-field (right side) distribution along the resonator structure for the resonator design producing the results shown in Fig. 2 in the main text. The E-field is concentrated along the parallel running wire sections, with its strength approaching zero along the narrow wire section. The opposite is the case for the H-field, where it is zero along the parallel wire sections and strongly concentrated along the narrow wire section. Note, that the H-field magnitude is homogeneous along the whole of the narrow wire section. The CST Microwave Studio at hand allowed us a simulation with perfect electric conductors. This is sufficient to model the general electric and magnetic energy distributions and resonance frequencies, however, not to simulate the corresponding oscillating magnetic field distribution, created by a superconducting rectangular wire. To this end, we numerically solve the Biot-Savart law for a rectangular wire cross-sectionS1, assuming a superconducting current distribution Jx;zS2, B1;x;z=m0 2pZw=2 w=2Zd=2 d=2Jr (x x0)2+ (z z0)2dx0dz0; (S1) with the vectors as J= (0;J(x;z);0)Tandr= (x x0;0;z z0)Tandm0being the magnetic constant. The integration is performed over the cross-section of the wire, of width wand thickness d. We define the wire cross-section in the x-z-plane, with win x- direction and din z-direction. The length of the wire is along the y-direction. For a superconducting wire, the current is not homogeneously distributed over the cross-section of the wire. Current is only flowing on the surface and is exponentially decaying towards the center of the wire. The characteristic length scale is given by the London penetration depth lL. We use the4 FIG. S3. Finite element simulations of resonator. CST Microwave Studio simulation of the distribution of E-fields and H-fields across the resonator structure. The color encoded fields represent the magnitude values. following expression for the current distributionS2 J(x;z) =J1 coshz0=lL coshd=lL" Ccoshx0=l1 coshw=l1+1 cosh x0=l2=coshw=l2p 1 (x0=w)2# +J2 J1coshx0=lL coshw=lL! ; (S2) where J2 J1=1:008 coshd=lLs w=l? 4l?=lL 0:08301lL=l?; C= 0:506p w=2l?0:75 ; l1=lLp 2lL=l?; l2=0:774 l2 L=l?+0:5152l?; l?=lL=2d: The prefactors J1andJ2define the amplitude of the current density and hence the absolute value of the oscillating magnetic field B1. We define J1by normalizing the vacuum B1field to the energy density stored in the resonatorS3,S4 1 2¯hwres 2=1 2m0Z B2 1dV=1 2m0B2 1Vm; (S3) with Vmrepresenting the resonator mode volume. The additional factor of1=2on the left hand side of S3 takes into account that only half of the total energy is stored in the magnetic fieldS5. As our resonator design is a quasi 1-dimensional structure we have to define boundaries for the mode volume in the x- and z-direction. A common assumption is to use the width of the conductor wire wS6. For simplicity, we approximate the x-z-area of the mode distribution with the area of an ellipse. For the last dimension we use the length of the narrow wire section, supported by the CST Microwave Studio simulations (see Fig. S3). In total we find5 the mode volume to be Vm= ((p3:0m2:025m) wd)300m=5696m3. Figure S4 shows the resulting distribution of the oscillating magnetic field for the cross-section of the rectangular wire of width w=2m and thickness d=50nm. The magnitudejB1;x;zjis encoded in the color and the arrows indicate the B1;xandB1;zcomponents of the oscillating field. -2000200z (nm)400 -400 x (μm)0 -1 1 2 3 -2 -320 15 10 5 |B| (nT) 30 25 FIG. S4. Cross-section of resonator magnetic field distribution. Calculated magnitude of the magnetic field distribution around the cross- section of a rectangular superconducting wire. The wire cross-section lies in the center, indicated by the grey rectangular. The red arrows show the direction of the magnetic field. With the simulated B1field distribution we can calculate the position dependent single photon - single spin coupling strength g0(r)S3,S4for each magnetic moment per unit cell of CGT (ab-plane 0 :68nmS7,S8, along the c-axis 0 :7nmS9). Summation over all CGT unit cells Nwithin the mode volume of the resonator results in the collective coupling strength geff=s N å i=1jg0(ri)j2=gCGTmB 2¯hs N å i=1jB1(ri)j2=gCGTmB 2¯hNys N å i=1h (B2 x;i+B2 z;i)i : (S4) Here, mBis the Bohr magneton, Nyis the number of unit cells along the y-direction and gCGTis the g-factor for CGT for which a value of 2 :18S10is used. Note, we give the collective coupling strength for spin1=2and for linear polarized microwavesS3. For the calculation of gefffor the resonator loaded with 15 monolayers of CGT we extracted its lateral dimension from the AFM measurements (see Fig. S4 (g)) to 2 m along the x-direction and 12 m along the y-direction. The flake is assumed to lie directly on top of the superconducting wire without any gap in between. For these values the simulation yields geff=2p=8:94MHz, which is about a factor 2 :5 larger than the experimentally determined value of 3 :61MHz. The overestimation of the simulation most likely results from non-ideal conditions in the experiment. The corresponding flake lies at the top end of the resonators narrow wire section (see Fig. S5 (g)), where B1is concentrated. The finite element simulations show that in this area the field strength is already declining, resulting in a reduced coupling strength. Further, AFM can overestimate the thickness of a flake slightly for when there is a gap between resonator surface and flakeS9. The calculation also not includes the multiple peaks observed in the experiment, which - depending on their real nature - can distribute the magnon density over all resonant peaks. Nevertheless, we can use the simulation to estimate the signal reduction by scaling down the thickness of the flake to a single monolayer. Reducing the simulation to a single monolayer, while keeping the lateral dimensions, results in geff=2p=2:33MHz, a reduction by a factor of 0 :26. IV. AFM MEASUREMENTS ON CGT FLAKES After the transfer of the CGT flakes onto the individual resonators and after measuring FMR, we characterized the thickness of the flakes by AFM. Figure S5 shows a selection of height profile maps from the three resonator chips, including a height profile along the inductor wire of the resonator (blue line in the AFM profile images in Fig. S5). To extract the thickness we fit the steps in the height profile (red or green lines in the height profiles in Fig. S5). Note, the height values are relative values with an arbitrary offset. Figure S5 (g) shows the thinnest flake of this study, where the processed FMR data is shown in Fig. 4 in the main text.6 10 μm 5 μm0 5 x (μm)y (nm)40506070 5 μmy (nm) 02040 0 4 x (μm)2 6 5 μmy (nm) 506070 0 4 x (μm)2 680 8 5 μm y (nm) 02040 0 10 x (μm)5 1560 2080 y (nm)050100 0 10 x (μm)5 5 μm 5 μm y (nm) 02040 0 10 x (μm)5 15 5 μm 10 μmy (nm)101520 0 6 x (μm)325 y (nm) 03060 0 20 x (μm)1090 y (nm)50100 0 6 x (μm)3150 120 30(c) (b) (a) (f) (e) (d) (i) (h) (g)9 FIG. S5. AFM measurements. AFM profile images with respective height profile (above) along the resonator inductor wire (blue and purple lines in profile images, with the arrow indicating scan direction). a-cfigures for resonator chip 1 (refer to Tab. I), having resonance frequencies with CGT of 17237MHz, 17790MHz and 16470MHz, respectively. d-ffigures for resonator chip 2 (refer to Tab. I), having resonance frequencies with CGT of 13244MHz, 12063MHz and 13620MHz, respectively. g-iimages for resonator chip 3 (refer to Tab. I), having resonance frequencies with CGT of 12314MHz, 13272MHz and 17054 ;MHz, respectively. The red and green solid lines are fits to the height profiles. V. ANALYSIS AND ADDITIONAL FMR DATA We analyze our experimental data, using the model functions (2) and (3) from the main text in a two-step semi-optimized fashion. The main intention for this approach is to minimize the number of free parameters in our model functions. In a first7 coarse step, we match the collective coupling strength geff;kto fit the experimental data, assume a constant separation between the individual magnon modes at BFMR ;kand the same magnon loss rate gfor all modes and determine the resonator loss rate k0from the resonator transmission far detuned from the FMR with the CGT flakes. This results in 3 free parameters for the first stage of our analysis, the magnon loss rate g,BFMR of the main mode and the constant separation between the BFMR ;k. After this first step we arrive at a best fit to the envelope of the experimental data, however with not matching amplitudes. In a consecutive second step, we manually optimize the geff;kto arrive at a model in good agreement with wresandkeff(see dashed lines in Fig. S6). Fig. S6 shows additional results from the corresponding FMR measurements performed on the in Fig. S5 showed resonators. As described in the main text, the measurements were performed at a temperature of 1 :8K and recording the microwave trans- missionjS21j2as a function of the static magnetic field. Analyzing the microwave transmission by fitting a Fano resonance lineshape to it we extract the effective loss rate of the resonator, interacting with the CGT keff. Figure S6 shows the resulting keffas a function of the magnetic field. In general, the response of the CGT FMR is complex and varies for the different res- onators. The resonance lineshape is not well described by just a single Lorentzian and requires multiple peaks to produce a good agreement. For some resonators, keffexhibits obvious peaks, residing on a broader spectrum (see Fig. S6 (c), (f) and (i)). Together with the observation of well and clearly separated peaks for the resonator loaded with the thinnest CGT flake of 11nm, we motivating the multiple peak analysis as presented in the main text. However, as the individual peaks are overlapping for the remainder of the resonators we only applied a qualitative analysis. κeff/2π (MHz)580 600 Magnetic Field (mT)62036912 525 550 Magnetic Field (mT)5751.61.82.0 600 625 675 700 Magnetic Field (mT)725246 7508 10 500 525 Magnetic Field (mT)5501.251.301.35 5751.40 520 560 Magnetic Field (mT)6000.900.951.00 6401.05 650 675 Magnetic Field (mT)7004812 72516 675 700 Magnetic Field (mT)7251.41.61.8 κeff/2π (MHz) 750 650 700 Magnetic Field (mT)750369 800 12 500 520 Magnetic Field (mT)540123 560 (c) (b) (a) (f) (e) (d) (i) (h) (g) κeff/2π (MHz) κeff/2π (MHz) κeff/2π (MHz) κeff/2π (MHz) κeff/2π (MHz) κeff/2π (MHz) κeff/2π (MHz) FIG. S6. Additional data on magnon-photon coupling of CGT-resonator devices. Results from FMR measurements with effective loss rate keff=2pas a function of the static magnetic field. a-cresults for resonator chip 1 (refer to Tab. I), having resonance frequencies with CGT of 17237MHz, 17790MHz and 16470MHz, respectively. d-fresults for resonator chip 2 (refer to Tab. I), having resonance frequencies with CGT of 13244MHz, 12063MHz and 13620MHz, respectively. g-iresults for resonator chip 3 (refer to Tab. I), having resonance frequencies with CGT of 12314MHz, 13272MHz and 11899MHz, respectively. The orange solid lines are semi-optimized fits, as described in the main text. The errorbars in the figures represent the standard deviation from the Fano resonance lineshape fit to the respective resonator transmission. Figure S7 shows the extracted collective coupling strength geffas a function of the square root of the FMR active volume. We define the active volume as the overlap of the oscillating magnetic field B1and the CGT flake lying on the resonator. The B1 field distribution, discussed in Sec. III, is used to estimate the extend of the B1and is taken as 2 m. From AFM measurements and microscope images we extract the thickness and lateral dimensions of the flakes to calculate the final active volume. As the collective coupling is proportional to the square root of the number of magnetic momentsS3, which are interacting with the resonator field, it follows that geffscales linearly with the square root of the active volume. This linear trend is highlighted by the orange solid line in Fig. S7. The majority of the extracted data follows this linear trend very well, corroborating our analysis. Only 3 data points deviate strongly from the rest of the data, which we attribute to significant inhomogeneities in the CGT-flakes, making the volume estimation inaccurate. These data points are highlighted in red in Fig. S7.8 0 1 2 3 4 51030 20Collective Coupling geff/2π (MHz) Square Root of Active FMR Volume ( μm3/2) FIG. S7. Scaling of the collective coupling. Collective coupling strength geffas a function of the FMR active CGT-flake volume. The orange line highlights the linear trend of geffwith increasing volume. The red symbols are regarded as outliers, as these flakes show inhomogeneities, leading to inaccurate volume estimations. The star symbol represents data from the thinnest flake (see data in Fig. 4 in the main text) and the pentagon symbol data from the 17nm flake (see data in Fig. 2 in the main text) The errorbars give confidence values for the extracted values. VI. MAGNETO-STATIC SPIN-WAVE DISPERSION IN THIN-FILM MAGNETS WITH PERPENDICULAR ANISOTROPY Here we describe the spin-wave mode frequency in a thin-film magnet with perpendicular anisotropy along the film normal. We consider this at the magnetic-dipole limit where the wavelength is relatively large and the exchange interaction contribution to the spin-wave dispersion is neglected. Furthermore, standing spin-wave modes along the thickness direction are also ruled out since these modes only appear at much higher frequencies than the main mode, where we consistently observe additional peaks at both higher and lower frequencies from the main mode. The mode (angular) frequency ( w) for wavevector k=0 when we apply a magnetic field Balong one of the film plane directions can be given by Eq. 3d in Ref.S11as: w g2 =B B+m0Ms 2Ku Ms : (S5) Here, g,MsandKuare the gyromagnetic ratio, saturation magnetization and the perpendicular anisotropy energy density, respec- tively. Note, that the total field within m0Ms 2Ku Msis negative for perpendicularly-magnetized films which we consider in this section. Within the magnetic-dipole limit, the demagnetization term m0Msis modified for spin-waves with finite k, depending on the relative orientation between the Msandkdirections. Here we follow the expression given in Serga et al.S12. For pure backward volume magnetostatic modes where kkMs(illustrated in Fig. S8), the mode frequency becomes: wBVMSW g2 =B B+m0Ms1 e kt kt 2Ku Ms ; (S6) where tis the thickness of the magnet. Note, that this expression is only valid for the case where Msis colinear to B, meaning that jBj>jm0Ms 2Ku Msj. To the limit of k!0, the term (1 e kt)=ktis reduced to unity, consistent to Eq. (S5). When kis nonzero, we can observe that wBVMSW becomes smaller than that for k= 0, exhibiting a negative group velocity for this spin-wave mode. As the opposite extreme where k?Ms(illustrated in Fig. S8), the resonance frequency becomes larger than that for k= 0 and is called magneto-static surface spin-wave mode. The mode frequency expression for this mode is given by: wMSSW g2 =B B+m0Ms 2Ku Ms +m2 0M2 s 1 e 2kt : (S7) Here, m2 0M2 s 1 e 2kt is the spin-wave correction term which goes to zero for k!0 (hence consistent to Eq. (S5)) and becomes positive for k>0, meaning that wMSSW becomes larger as soon as spin-waves gain momentum along this direction. We use these two expressions in an effort to explain the origin of the multiple peaks in our experiments. Figure S8 plots the calculated wBVMSW =2pandwMSSW =2pas a function of wavevector k. The range of wavevector is chosen such that the resulting resonance frequencies are within the same order of magnitude as the observed mode splittings in the experiment ( µ100MHz).9 Wavevector (μm-1)6 4 8 10 12 2 012.712.812.9Resonance Frequency (GHz)13.0 12.6 12.5 12.4Kittel BVMSW MSSWB0B0100 MHz FIG. S8. Spin-wave dispersion. Spin-wave resonance frequency for BVMSW (green solid line) and MSSW (yellow solid line) as a function of wavevector. The dashed blue line is the resonance frequency of the k=0 main mode. The parameters used are B0=598mT, gCGT =2:18, m0Ms=194:3mT and Ku=3:84104J=m3and a thickness of 17nm. The grey area highlights a 100MHz margin relative to the main mode, indicating the order of magnitude of the mode splitting observed in the experiment. The arrows on the right hand side illustrate the relative wavevector orientations of the BVMSW and MSSW spin-wave modes with respect to the static magnetic field. The corresponding wavelength to a 100MHz resonance offset to the main mode are about 2 :2m and 620nm for wBVMSW andwMSSW , respectively. These values are within a reasonable scale for our different lateral CGT flake dimensions under investigation. This suggests that spin-wave modes are likely the origin of the multiple resonance peaks observed. The thinnest CGT flake shows, however, a deviation from this behaviour. We only observe modes at lower frequencies, which would indicate to BVMSW modes. Calculating the respective shortest wavelength results in 225nm, which is significantly shorter than for the other devices. We assume that the placement and irregular shape are likely to cause this difference. First, this flake is placed at the very edge of the inductor wire, where the B1field strength is declining (see Fig. S3), reducing the FMR active area. Thickness steps can lead to a wavelength down-conversionS13, however, with the overall irregular shape of the flake it is difficult to define a length scale for a standing spin wave mode. VII. ATOMISTIC SPIN DYNAMICS SIMULATIONS OF FMR To study the ferromagnetic resonance in CGT we perform atomistic spin dynamics simulationsS14,S15. The magnetic Hamil- tonian employed in the simulations is given by: H= 1 2å i;jSiJi jSj å iDi(Si·e)2 å imiSi·(B0+B1) (S8) where i,jrepresent the atoms index, Ji jrepresents the exchange interaction tensor, Dithe uniaxial anisotropy, which for CGT is orientated out of plane ( e= (0;0;1)) and B0the external static magnetic field applied in-plane during the ferromagnetic resonance simulations and B1=B1sin(2pnt)the oscillating field applied perpendicular with respect to B0. The CGT system has been parameterized from first principle methodsS9, up to the third nearest neighbor intralayer and interlayer exchange. The exchange values have also been re-scaled by Gong et al.S9with a 0.72 factor to obtain the experimental TCand multiplied by S2to match the magnetic Hamiltonian. The magnetic moment or Cr is considered 3.26 mBS16and the uniaxial anisotropy has a value of 0 :05 meV as extracted from first principle methodsS9. The parameters used in the simulations are given in Table II. FMR calculations have previously been employed for atomistic models, and can reproduced well the variation of linewidth with temperature, for example, in recording media systemsS17. Hence, in the current simulations we use the same setup of frequency swept FMRS17and we obtain the spectra by performing a Fourier transform of the magnetisation component parallel to the oscillating field. Since these calculations are done close to 0K, no averaging is require to reduce the thermal noise. To excite the FMR mode, we apply a DC field in-plane of 0.9 T on x-direction and an AC field perpendicular to the DC field, on y-direction. The Fourier transform has been performed for the y-component of magnetisation for 5ns after an initial 1ns equilibration time. A thermal bath coupling has been chosen in agreement with the upper limit of the Gilbert damping observed in experiments. The system size we performed FMR on is a 4-layer CGT system, with lateral size of 6 :91nm11:97nm, periodic boundary conditions in xy and total of 1600 atoms. The small system size has been used to reduce the computational cost associated10 Quantity Symbol quantity units Timestep ts 0.1 fs Thermal bath coupling a 0.02 Gyromagnetic ratio ge 1.7608591011rad s 1T 1 Magnetic moment mB 3.26S16mB Uniaxial anisotropy Di 0.05S9meV/link Simulation temperature T 0.001 K Static magnetic field B0 0.9, 0.7 T Oscillating magnetic field amplitude B0 0.001 T FMR frequency n varied GHz Intralayer exchange, NN J1 2.71S9meV/link Intralayer exchange, 2NN J2 - 0.058S9meV/link Intralayer exchange, 3NN J3 0.115S9meV/link Interlayer exchange, NN Jz 1-0.036S9meV/link Interlayer exchange, 2NN Jz 20.086S9meV/link Interlayer exchange, 3NN Jz 30.27S9meV/link TABLE II. Simulation parameters for FMR on CGT system . with FMR simulations. Experiments have showed modified g-factors due to photon-magnon coupling hence hereby we propose a simple model where the properties of the individual layers have been modified to include different gyromagnetic ratio, as illustrated in Fig. S9 a. We can define the resonance frequencies for each magnetic layer using the Kittel equation in the case of in-plane applied field with perpendicular anisotropy B?u: w=gp B0(B0 B?u) (S9) We next investigate the FMR signal for a few cases assuming the CGT monolayers at low or strong interlayer exchange cou- plings J0 z=0;0:1%;10%;100% Jz, where Jzcorresponds to the pristine interlayer exchange (Fig. S9 b-c). In the low interlayer exchange regime ( J0 z=0;0:1%Jz), the CGT presents multiple peaks with each frequency corresponding to the layer dependent gyromagnetic ratio, g-n(g1) =16:81GHz, n(g2) =25:22GHz, n(g3) =33:62GHz. At J0 z=0:1%J0 z(Fig. S9 b) we can still observe resonance peaks corresponding to each individual layer. However by increasing the exchange coupling to 10% J0 zor higher (Fig. S9 c) there is a single FMR peak indicating that the system behave coherently with all layers having the same FMR frequency. The single FMR frequency corresponds to the average magnetic properties of the CGT layers. Small variations of the resonance frequency as function of the inter-layer exchange coupling can be observed which these being correlated to the transition of the system from the multi-peaks regime to a coherent excitation. By calculating the damping of the highest reso- nance peaks from a Lorenzian fit, we reobtain the damping corresponding to the input thermal bath coupling, 0 :02 with a relative tinny error5%. Overall, the interlayer exchange coupling locks the dynamics of individual layers coherently together without allowing multiple frequencies at the FMR signalS18. [S1]A. E. Primenko, M. A. Osipov, and I. A. Rudnev, Technical Physics 62, 1346 (2017). [S2]L. H. Lee, T. P. Orlando, and W. G. Lyons, IEEE Transactions on Applied Superconductivity 4, 41 (1994). [S3]C. W. Zollitsch, K. Mueller, D. P. Franke, S. T. B. Goennenwein, M. S. Brandt, R. Gross, and H. Huebl, Applied Physics Letters 107, 142105 (2015). [S4]S. Weichselbaumer, P. Natzkin, C. W. Zollitsch, M. Weiler, R. Gross, and H. Huebl, Physical Review Applied 12, 024021 (2019). [S5]R. J. Schoelkopf and S. M. Girvin, Nature 451, 664 (2008). [S6]D. I. Schuster, A. A. Houck, J. A. Schreier, A. Wallraff, J. M. Gambetta, A. Blais, L. Frunzio, J. Majer, B. Johnson, M. H. Devoret, S. M. Girvin, and R. J. Schoelkopf, Nature 445, 515 (2007). [S7]Y . F. Li, W. Wang, W. Guo, C. Y . Gu, H. Y . Sun, L. He, J. Zhou, Z. B. Gu, Y . F. Nie, and X. Q. Pan, Physical Review B 98, 125127 (2018). [S8]Y . Sun, R. C. Xiao, G. T. Lin, R. R. Zhang, L. S. Ling, Z. W. Ma, X. Luo, W. J. Lu, Y . P. Sun, and Z. G. Sheng, Applied Physics Letters 112, 072409 (2018). [S9]C. Gong, L. Li, Z. Li, H. Ji, A. Stern, Y . Xia, T. Cao, W. Bao, C. Wang, Y . Wang, Z. Q. Qiu, R. J. Cava, S. G. Louie, J. Xia, and X. Zhang, Nature 546, 265 (2017). [S10]S. Khan, C. W. Zollitsch, D. M. Arroo, H. Cheng, I. Verzhbitskiy, A. Sud, Y . P. Feng, G. Eda, and H. Kurebayashi, Physical Review B 100, 134437 (2019). [S11]M. Farle, Reports on Progress in Physics 61, 755 (1998). [S12]A. A. Serga, A. V . Chumak, and B. Hillebrands, Journal of Physics D: Applied Physics 43, 264002 (2010). [S13]J. Stigloher, T. Taniguchi, M. Madami, M. Decker, H. S. Körner, T. Moriyama, G. Gubbiotti, T. Ono, and C. H. Back, Applied Physics Express 11, 053002 (2018). [S14]D. A. Wahab, M. Augustin, S. M. Valero, W. Kuang, S. Jenkins, E. Coronado, I. V . Grigorieva, I. J. Vera-Marun, E. Navarro-Moratalla, R. F. Evans, et al. , Advanced Materials 33, 2004138 (2021). [S15]A. Kartsev, M. Augustin, R. F. Evans, K. S. Novoselov, and E. J. G. Santos, npj Computational Materials 6, 1 (2020). [S16]I. A. Verzhbitskiy, H. Kurebayashi, H. Cheng, J. Zhou, S. Khan, Y . P. Feng, and G. Eda, Nature Electronics 3, 460 (2020).11 FIG. S9. Atomistic simulations. a, Schematic of the crystal structure of CGT with atoms defined by different colours. b,FMR spectra of 4 layer CGT where the layers are low interayer exchange coupled (0 ;0:1%J0z, where J0zis the pristine CGT interlayer exchange). c,Similar as b, but with the layers at a strong exchange coupling (10% ;100% J0z). The solid lines in b-crepresent a Lorenzian fit to the numerical data. [S17]M. Strungaru, S. Ruta, R. F. Evans, and R. W. Chantrell, Physical Review Applied 14, 014077 (2020). [S18]Data inputs/plots utilised for Supplementary Figure S7 (atomistic simulations) can be found at the following GitHub repository. | 2022-06-06 | Layered van der Waals (vdW) magnets can maintain a magnetic order even down
to the single-layer regime and hold promise for integrated spintronic devices.
While the magnetic ground state of vdW magnets was extensively studied, key
parameters of spin dynamics, like the Gilbert damping, crucial for designing
ultra-fast spintronic devices, remains largely unexplored. Despite recent
studies by optical excitation and detection, achieving spin wave control with
microwaves is highly desirable, as modern integrated information technologies
predominantly are operated with these. The intrinsically small numbers of
spins, however, poses a major challenge to this.
Here, we present a hybrid approach to detect spin dynamics mediated by
photon-magnon coupling between high-Q superconducting resonators and ultra-thin
flakes of Cr$_2$Ge$_2$Te$_6$ (CGT) as thin as 11\,nm. We test and benchmark our
technique with 23 individual CGT flakes and extract an upper limit for the
Gilbert damping parameter. These results are crucial in designing on-chip
integrated circuits using vdW magnets and offer prospects for probing spin
dynamics of monolayer vdW magnets. | Probing spin dynamics of ultra-thin van der Waals magnets via photon-magnon coupling | 2206.02460v2 |
Evaluating Gilbert Damping in Magnetic Insulators from First Principles Liangliang Hong,1, 2Changsong Xu,1, 2and Hongjun Xiang1, 2,∗ 1Key Laboratory of Computational Physical Sciences (Ministry of Education), Institute of Computational Physical Sciences, State Key Laboratory of Surface Physics, and Department of Physics, Fudan University, Shanghai 200433, China 2Shanghai Qi Zhi Institute, Shanghai 200030, China (Dated: September 21, 2023) Magnetic damping has a significant impact on the performance of various magnetic and spin- tronic devices, making it a long-standing focus of research. The strength of magnetic damping is usually quantified by the Gilbert damping constant in the Landau-Lifshitz-Gilbert equation. Here we propose a first-principles based approach to evaluate the Gilbert damping constant contributed by spin-lattice coupling in magnetic insulators. The approach involves effective Hamiltonian mod- els and spin-lattice dynamics simulations. As a case study, we applied our method to Y 3Fe5O12, MnFe 2O4and Cr 2O3. Their damping constants were calculated to be 0 .8×10−4, 0.2×10−4, 2.2×10−4, respectively at a low temperature. The results for Y 3Fe5O12and Cr 2O3are in good agreement with experimental measurements, while the discrepancy in MnFe 2O4can be attributed to the inhomogeneity and small band gap in real samples. The stronger damping observed in Cr 2O3, compared to Y 3Fe5O12, essentially results from its stronger spin-lattice coupling. In addition, we confirmed a proportional relationship between damping constants and the temperature difference of subsystems, which had been reported in previous studies. These successful applications suggest that our approach serves as a promising candidate for estimating the Gilbert damping constant in magnetic insulators. I. INTRODUCTION Recent decades have witnessed rapid developments in magnetics and spintronics [1–3]. A long-time pursuit in spintronics is to actively control and manipulate the spin degrees of freedom in solid-state systems. Related fun- damental studies involve spin transport, spin dynamics and spin relaxation [4]. Within these domains, magnetic damping often plays a crucial role. Generally, stronger damping enables a faster writing rate for magnetic mem- ories, while lower damping leads to a longer propagation distance of spin waves. Therefore, it is always essential to accurately evaluate the magnetic damping in different materials. For instance, yttrium iron garnet (YIG) is a highly promising spintronic material due to its ultra-low magnetic damping [5–7]. However, the intrinsic mecha- nism behind its unique property has yet to be fully eluci- dated, which partly motivates us to carry out this work. At present, magnetic damping is typically represented by a phenomenological term in the well-known Landau- Lifshitz-Gilbert (LLG) equation, which has been widely employed to simulate magnetization dynamics [8, 9]. A basic form of this equation can be written as, ∂ ⃗ m ∂t=−γ ⃗ m×⃗B+α m⃗ m×∂ ⃗ m ∂t(1) where ⃗Brepresents the total magnetic field acting on the local dipole ⃗ m,mdenotes the norm of ⃗ m,γis the gyro- magnetic ratio, and αis the Gilbert damping constant. The second term on the right side, as we mentioned, leads ∗hxiang@fudan.edu.cndirectly to the relaxation process, in which the rate of en- ergy dissipation is determined by the damping constant. Given the importance of αin magnetization dynamics, its origin has been extensively studied in the literature [10–13]. To the best of our knowledge, both intrinsic and extrinsic mechanisms contribute to the damping. Specif- ically, the intrinsic factors include spin-lattice and spin- electron couplings, while the extrinsic contributions pri- marily involve lattice imperfections and eddy currents [14, 15]. Two types of first-principles based methods have been developed to calculate the damping constants in the past. One approach involves the breathing Fermi surface model [16, 17] and the torque correlation model [18, 19], while the other is based on the scattering theory from linear response [20–22]. These methods have demonstrated re- markable success in studying the magnetic damping in transition metals such as Fe, Co, and Ni. Despite be- ing free from complicated experiments, which are mostly based on ferromagnetic resonance, these theoretical ap- proaches still exhibit several limitations. Firstly, when dealing with complex systems, we often have to spend a significant amount of computing resources on the first- principles calculations. In addition, these methods are more suitable for calculating the electronic contribution to Gilbert damping in metallic magnets, thus rarely tak- ing the effect of spin-lattice coupling into consideration [14, 23]. Recently, spin-lattice dynamics (SLD) simulations [24] have been adopted as an alternative method to evaluate the Gilbert damping parameters. In Ref. [23], the au- thors constructed an empirically parameterized Hamil- tonian model for a cobalt cluster. They coupled a pre- heated lattice with a fully ordered spin state, then per- formed SLD simulation. During the relaxation process,arXiv:2309.11152v1 [cond-mat.mtrl-sci] 20 Sep 20232 the energy of lattice and spin subsystems were recorded and fitted to the following logistic functions, Ulat=Ulat 0−∆U0 1 + exp[ −η∆U0t−Θ](2) Umag=Umag 0+∆U0 1 + exp[ −η∆U0t−Θ](3) from which they extracted the relaxation rate Γ = η∆U0 and calculated the damping constant α=ηµS/γ. Here, µSdenotes the magnitude of magnetic moments. In Ref. [25], the authors also built an empirical potential model for a periodic bcc Fe system. They firstly applied an ex- ternal magnetic field in the z-direction and thermalized the system to a finite temperature. Then, the magnetiza- tion orientation of each atom was rotated artificially by a same angle. Afterwards, the system would relax back to equilibrium, during which the averaged z component of atomic magnetization was recorded and fitted to the following function, mz(t) = tanhα 1 +α2γBext(t+t0) (4) where αwas exactly the Gilbert damping parameter to be estimated. Since these works selected transition met- als as the research object, their results were both orders of magnitude smaller than the experimental values. In addition, the use of empirically parameterized models re- duced the accuracy of their simulated results. In this work, we combine SLD simulations with first- principles based effective Hamiltonian models to evalu- ate the damping constants in magnetic insulators, where the dominant contribution results from spin-lattice cou- plings. Compared to the previous studies, our work has made improvements mainly in two aspects. Firstly, the utilization of first-principles based Hamiltonian models in simulations enhances the reliability of our conclusions. Besides, the better choice of research objects allows for demonstrating the superiority of SLD simulations. In particular, the microscopic origin of low damping in YIG will be investigated. The paper is organized as follows. In Sec. II, we introduce our effective Hamiltonian model, parameterization methods, and a scheme for evaluating Gilbert damping parameters. Then, both the validation and application of our method are presented in Sec. III. Finally, we summarize this work and give a brief outlook in Sec. IV. II. MODEL AND METHODS This section is split into three parts. Firstly (in Sec. II A), we introduce a generic form of our effective Hamil- tonian model. Then, methods involving the calculation of model parameters are presented in Sec. II B. At the last part (Sec. II C), we propose a novel scheme to de- termine the Gilbert damping constant through dynamics simulations.A. The Hamiltonian Model Since our purpose is to evaluate the contribution of spin-lattice coupling to magnetic damping, the effective Hamiltonian model must incorporate both spin and lat- tice degrees of freedom. A concise and generic formula that meets our basic requirements consists of the three terms as follows: H=HL({ui,α}) +HS({⃗ sj}) +HSLC({ui,α,⃗ sj}) (5) where αabbreviates three orthogonal axes, ui,αrepre- sents the displacement of atom i, and ⃗ sjis a unit vector that represents the direction of spin j. The first term HLin Hamiltonian model describes the dynamical behavior of individual phonons. Technically, we take the atomic displacements as independent vari- ables and expand the Hamiltonian to the second order with Taylor series. Then, we have the form as, HL=1 2X ijX αβKij,αβui,αuj,β+1 2X i,αMi˙ui,α˙ui,α(6) where Kij,αβ denotes the force constant tensor and Mi represents the mass of atom i. Similarly, the second term HSdescribes the dynami- cal behavior of individual magnons. For simplicity but no loss of accuracy, we only considered the Heisenberg exchange interactions between neighbor magnetic atoms in this work, though more complex interactions could be taken into account in principle. Therefore, this term can be expressed as, HS=X ⟨i,j⟩Jij⃗Si·⃗Sj (7) where Jijdenotes the isotropic magnetic interaction co- efficient. The third term HSLCrepresents the coupling of spin and lattice subsystems, and is expected to describe the scattering process between phonons and magnons. As an approximation of the lowest order, this term can be written as, HSLC=X ⟨i,j⟩X kα∂Jij ∂uk,αuk,α ⃗Si·⃗Sj (8) According to the theory of quantum mechanics, this coupling term provides a fundamental description of the single-phonon scattering process, which is believed to be dominant among all scatterings in the low-temperature region. This type of relaxation mechanism in ferromag- netic resonance was systematically studied by Kasuya and LeCraw for the first time [26]. It’s worth noting that a higher order of Taylor expansion could have been con- ducted to improve the accuracy of Hamiltonian models directly. For instance, the scattering between individual phonons can be adequately described by the anharmonic terms. However, as one always has to make a trade-off3 between the precision and complexity of models, in this work we choose to neglect the high order terms since the anharmonic effects in current investigated systems are not important. In this study, we adopted the symmetry-adapted clus- ter expansion method implemented in the Property Anal- ysis and Simulation Package for Materials (PASP) [27] to build the Hamiltonian model presented above. This package can identify the nonequivalent interactions and equivalent atom clusters in a crystal system by analyz- ing its structural properties based on the group theory. A significant benefit of working with PASP is we are en- abled to describe the target system with the least number of parameters. In the next section, we will discuss how to calculate the model parameters for different materials. B. Calculation of Model Parameters Firstly, the Heisenberg exchange coefficients Jijand spin-lattice coupling constants ∂Jij/∂uk,αcan be calcu- lated with the four-state method [28, 29]. The basic flow is to construct four artificially designated spin states of the target system, calculate the corresponding energies and forces based on the density functional theory (DFT), then determine the parameters by proper combination of those results. At the last step, the following formulas will be used, Jij=E↑↑+E↓↓−E↑↓−E↓↑ 4S2(9) ∂Jij ∂uk,α=F↑↑ k,α+F↓↓ k,α−F↑↓ k,α−F↓↑ k,α 4S2(10) where Sis the spin quantum number of magnetic atoms, Eis the total energy of system and Fk,αrefers to one component of the force on atom k. The superscripts ( ↑↑, ↓↓,↑↓,↓↑) specify the constrained spin states of system in the calculation. More technical information about the four-state method can be found in the references [28, 29]. Compared to other approaches, the four-state method of- fers an obvious advantage in that no additional DFT cal- culations are needed to determine the coupling constants ∂Jij/∂uk,αonce the exchange coefficients Jijhave been obtained. This is because the energy and forces are typ- ically provided simultaneously by one DFT calculation. Since atomic masses Mican be directly obtained from the periodic table, more efforts are needed to deal with the force constant tensor Kij,αβ. Currently, there are two commonly adopted ways to calculate the force constant tensor: density functional perturbation theory (DFPT) and finite displacement method. Both of these methods are applicable to our task. However, we cannot directly take the force constant tensor obtained from first-principles calculations as the model parameter. This is because in dynamics simula- tions we usually expand crystal cells to reduce the un- desired influence of thermal fluctuations, which leads toa conflict between the periodic boundary condition and the locality (also known as nearsightedness [30, 31]) of models. To be more specific, when calculating the con- tribution of one atom or spin to the total energy, we tend to set a well designed cutoff radius and ignore the inter- actions beyond it. This step is essential when dealing with a large-scale system, otherwise we will suffer from the model complexity and the computational cost. Nev- ertheless, if we set the elements of Kij,αβ that represent out-of-range interactions to be zero and leave the others unchanged, we may violate the so-called acoustic sum- mation rules: X iKij,αβ = 0 for all j, α, β. (11) It should be pointed out that a straightforward en- forcement of the acoustic summation rules, achieved by subtracting errors uniformly from force constants, will break the inherent crystal symmetry inevitably, which is the technique employed in phonopy [32]. To address the above issues, we adopted a more appropriate method in this work. Before a detailed introduction, it’s necessary to recall that not every element of the force constant ten- sor serves as an independent variable due to the crystal symmetries. Taking the cubic cell of Y 3Fe5O12(contain- ing 160 atoms) for example, there are 230400 elements in the tensor. After symmetry analyses, we find that only 597 independent variables {pn}are needed to adequately determine all the tensor elements {Kij,αβ({pn})}, where the effect of locality is already considered. Afterwards, our method is to set a correction factor xnfor each vari- ablepnand minimize the deviation of parameters under the constraints of Eq. (11). A mathematical reformula- tion of this method can be written as, min {xn}X n(xn−1)2,with X iKij,αβ({xnpn}) = 0 for all j, α, β.(12) In the case of Y 3Fe5O12, there are only 18 linearly inde- pendent constraints, which allow the extremum problem to be solved rigorously. The modified force constant ten- sor restores positive definiteness and translational sym- metry while maintaining the crystal symmetries. There- fore, the modified tensor meets the requirements for dy- namics simulations. In Sec. III B, the effectiveness of this approximate method will be demonstrated through a spe- cific example. All the first-principles calculations mentioned in this section are carried out using the Vienna ab initial simu- lation package (VASP) [33–35]. The force constants and phonon spectra are obtained by phonopy [32]. The opti- mizations formulated in (12) are accomplished with the function optimize.minimize implemented in SciPy [36].4 C. Evaluation of Damping Constants After the construction and parameterization of Hamil- tonian models, we are finally able to perform spin-lattice dynamics simulations. Before the evaluation of Gilbert damping constants, we briefly introduce the framework of SLD to cover some relevant concepts. In practice, the motion of magnetic moments follows the stochastic Lan- dau–Lifshitz–Gilbert (SLLG) equation [14], d⃗ mi dt=−γL⃗ mi× ⃗Bi+⃗Bfl i −γLα⃗ mi |⃗ mi|×h ⃗ mi× ⃗Bi+⃗Bfl ii (13) where γLis the renormalized gyromagnetic ratio, ⃗Bi= −∂H/∂ ⃗ m iis the effective local magnetic field and ⃗Bfl i refers to a stochastic field introduced by Langevin ther- mostat. At the same time, the motion of atoms obeys the Newton’s equation, d˙ui,α dt=1 Mi ⃗Fi,α+⃗Ffl i,α −ν˙ui,α (14) where νis the damping constant and ⃗Ffl i,αrefers to a stochastic force caused by thermal fluctuations. In this work, ⃗Bfl iand⃗Ffl i,αare modeled as normally distributed noises with temperature-dependent variances, Bfl i,β∼N 0,p 2αkBTS/γ|⃗ mi|δt (15) Ffl i,β∼N 0,p 2νMikBTL/δt (16) where TSandTLrefer to the equilibrium temperature of spin and lattice subsystems respectively. During simula- tions, we can also measure the transient temperature of each subsystem with the following formulas [37], TS=P i|⃗ mi×⃗Bi|2 2kBP i⃗ mi·⃗Bi, TL=1 2kBNX i,αMi˙u2 i,α (17) In this work, the LLG equation is numerically solved with the semi-implicit SIB method proposed by Mentink et al. [38]. The Newton’s motion equation is integrated using the Grønbech-Jensen-Farago Verlet-type method [39]. To ensure the stability of those algorithms, a step length of 0 .5 or 0 .2 fs is adopted [40], where the shorter one is used in energy-conserving simulations. Based on the combination of atomistic spin dynamics (ASD) and SLD simulations, a new scheme is proposed to evaluate the damping constant in magnetic materials. Here is the basic flow of this method and more details of a specific application are presented in Sec. III B. 1. Freeze the spin degree of freedom and thermalize the lattice from 0 to TLin the simulation. 2. Fix atomic positions and raise the temperature of spin to TS> TL. Compared to TL> TS, this type of nonequilibrium state is more common in actual scenarios.3. Perform an energy-conserving SLD simulation to relax the system. Normally, the spin temperature will decrease to the same as lattice and stay there till the end. 4. Conduct a series of ASD simulations with different Gilbert damping constants. The initial states are the same as in step 3 and the equilibrium temper- atures are set to be TL. 5. Compare the cooling rates ∂TS/∂tof spin system between SLD and ASD simulations to evaluate the equivalent Gilbert damping constant contributed by spin-lattice coupling. The key point behind step 5 is that the cooling rates observed in ASD simulations are related to the assigned damping constants, while in SLD simulation the cooling rate is determined by the strength of spin-lattice cou- pling. Note that the former relation can be viewed as a natural deduction of the LLG equation, ∂TS ∂t=1 CV∂Emag ∂t∝ −1 CV∂ ⃗ m ∂t·⃗B ∝ −1 CVα m⃗ m×∂ ⃗ m ∂t ·⃗B ∝α (18) where we have used Eq. (1) and simplified the formula of magnetic energy as Emag∝ −⃗ m·⃗B. III. RESULTS This section is divided into four parts. In Sec. III A, several test results are presented to validate the accu- racy of SLD simulations, which are implemented in the PASP package. Subsequently, detailed calculations on three magnetic materials, namely Y 3Fe5O12, MnFe 2O4 and Cr 2O3, are discussed in the rest parts. A. Validations In order to guarantee the reliability of our conclusions obtained from dynamics simulations, a series of pretests were carried out. We select some representative results and present them in Fig. 1, where Cr 2O3is taken as the object to be studied. Firstly, we set the ground state of Cr 2O3as the ini- tial state and performed a NVT simulation with Tset= 150K. As shown in Fig. 1(a), the temperature of spin and lattice subsystems increased to 150 Kin less than 5 ps and stayed there till the end. Since we can approxi- mate Ek= 0.5ELandEp= 0.5EL+ES, Fig. 1(b) also indicates that the contribution of phonons and magnons to the excited state energy is around 87.5% and 12.5% respectively. This result could be verified from another perspective. Note that there are totally 10 atoms in the5 FIG. 1. NVT and NVE relaxations of a spin-lattice coupled system (Cr 2O3) within the framework of spin-lattice dynamics. The top row plots the time evolution of temperatures and the bottom row shows the variation of potential, kinetic and total energies. (a) & (b): NVT thermalization from TL=TS= 0KtoTL=TS= 150 K. (c) & (d): NVE relaxation with TL= 30K, TS= 175 Kinitially. (e) & (f): NVE relaxation with TL= 180 K,TS= 30Kinitially. unit cell of Cr 2O3, which contribute 30 kBto the heat ca- pacity. Meanwhile, the 4 magnetic atoms will contribute another 4 kBin the low temperature region. Therefore, we can estimate that the contribution of magnons to the total heat capacity is close to 11.8%, which is consistent with the result from dynamics simulations. In Figs. 1(c) & 1(d), the initial state was set to be a nonequilibrium state with TL= 30KandTS= 175 K. As we expected, the total energy was well conserved when the system evolved to equilibrium. In addition, the final temperature fell within the range of 48 K∼55K, which agrees with our previous analysis of the heat capacities. Lastly, we simulated the relaxation process using an- other nonequilibrium excited state with TL= 180 Kand TS= 30Kas the initial state. As shown in Figs. 1(e) & 1(f), the temperature of spin system increased gradually to equilibrium with the total energy conserved through- out the simulation. Also, the final temperature is around 160K, which matches well with our analysis. It should be pointed out that there exist two notable differences be- tween this case and the previous. Firstly, the subsystems ultimately evolved to a same temperature in a finite time,which alleviated our concerns about the accuracy of SLD simulations. Besides, the relaxation time ( τ2) was much longer than that ( τ1) in Fig. 1(c). For this phenomenon, a qualitative explanation is presented below. Based on the theory of second quantization, the Hamil- tonian model presented in Sec. II A can be expressed in the following form [41, 42], HL=X qpℏωqp(b† qpbqp+ 1/2) (19) HS=X λϵλa† λaλ+Const. (20) HSLC=X λ,qpMλ,qpa† λ−qaλ b† qp−b−qp (21) where bqpdenotes the annihilation operator of phonons with wave vector qin branch p, and aλrepresents the an- nihilation operator of magnons with wave vector λ. All the parameters, namely ωqp,ϵλandMλ,qp, can be deter- mined from the effective Hamiltonian model in principle. According to the Fermi’s golden rule, we have W{nλ−q, nλ, Nqp→nλ−q+ 1, nλ−1, Nqp+ 1}=2π ℏ|Mλ,qp|2(nλ−q+ 1)( nλ)(Nqp+ 1)δ(ϵλ−q−ϵλ+ℏωqp) (22) W{nλ−q, nλ, N−qp→nλ−q+ 1, nλ−1, N−qp−1}=2π ℏ|Mλ,qp|2(nλ−q+ 1)( nλ)(N−qp)δ(ϵλ−q−ϵλ−ℏω−qp) (23)6 FIG. 2. (a) The primitive cell of Y 3Fe5O12. The golden balls represent iron atoms, the cyan ball stand for yttrium atoms, and the red balls refer to oxygen atoms. (b) The magnetic ground state of YIG. The arrows of different colors represent the spin directions of Fe atoms. (c) The density of states ob- tained by DFT calculations. (d) The temperature dependence of average magnetization measured in MC and ASD simula- tions. For YIG, the phase transition point from ferrimagnetic to paramagnetic lies in 530 K approximately. where Wrepresents the probability of one-phonon emis- sion or absorption, nλdenotes the occupation number of magnons and Nqpstands for phonons. Both nλandNqp can be evaluated approximately using the Bose–Einstein distribution. According to the above formulas, the scat- tering rate Wgrows linearly with Nand quadratically with n. Compared to Fig. 1(c), there are more phonons but fewer magnons in the case of Fig. 1(e), thus leading to a lower transition probability and a longer relaxation time. More technical details about the second quantiza- tion of interactions between phonons and magnons can be found in Ref. [41, 42]. B. Damping constants in Y 3Fe5O12 In the field of spintronics, Y 3Fe5O12(yttrium iron gar- net, YIG) has gained much attention due to its ultra-low magnetic damping [5–7]. The unique property of this material motivated us to investigate the intrinsic mecha- nism behind it. The crystal structure of YIG is presented in Fig. 3(a). There are totally 80 atoms in the primitive cell, of which 12 Fe ions are located in the center of oxy- gen tetrahedrons while the other 8 Fe ions are sited in oxygen octahedrons. The magnetic ground state of YIG is illustrated in Fig. 3(b). The Fe ions situated in differ- ent chemical environments contribute spins in opposite directions, which makes YIG a typical ferrimagnetic ma- terial.TABLE I. The Heisenberg exchange coefficients J of YIG, where an effective spin S= 1 is adopted. For the FeO−FeO pairs, the Greek letters ( α&β) refer to different chemical environments. All the results are calculated with the four- state method. Spin Pair. Distance (Angst) J (meV) 1NN FeT−FeO3.445 47.414 1NN FeT−FeT3.774 2.399 1NN FeO−FeO(α) 5.337 0.538 1NN FeO−FeO(β) 5.337 5.055 2NN FeT−FeO5.555 0.285 2NN FeT−FeT5.765 3.437 In order to evaluate the Gilbert damping constants in YIG, our first step is to prepare an effective Hamilto- nian model. Considering the balance between precision and efficiency, the cutoff radius of interactions was set to be 11.0 Bohr for atomic pairs and 6.7 Bohr for 3- body clusters. After symmetry analyses, we identified 612 nonequivalent interactions in total, which included 6 Heisenberg exchange terms and 9 spin-lattice coupling terms. To determine the interaction parameters, we carried out a series of first-principles calculations, where a cu- bic cell was adopted to reduce the interference between adjacent cells caused by periodic boundary conditions. Following the settings in Ref. [43], we utilized the pro- jector augmented-wave (PAW) method [44] and revised Perdew-Burke-Ernzerhof exchange-correlation functional for solids (PBEsol) [45] in our calculations. Besides, the DFT+U method in its simplified form [46] was employed where the effective Hubbard U parameter was set to be 4 eV for the 3 delectrons of Fe ions. In addition, a cutoff energy of 520 eV for plane wave basis and a Γ-centered 2×2×2 mesh of k-points were used in the DFT calcu- lations. In Figure 2(c), we present the density of states (DOS) for YIG. With a band gap of 1.863 eV, there is hardly any electric current occurring in the low temperature re- gion. Moreover, the Heisenberg exchange coefficients of YIG is listed in Table I. To verify the accuracy of these parameters, we conducted both Monte Carlo (MC) and ASD simulations. The temperature dependence of aver- age magnetization is shown in Fig. 2(d), which reveals the critical temperature of YIG to be 530 K. This result is slightly lower than the measured Curie temperature, TC= 560 K[5], but falls within our tolerance. The cal- culated results of coupling constants are provided in the supplementary material. Next, we come to deal with the force constant tensor. In order to demonstrate the impact of locality and val- idate the effectiveness of our optimization method, we present some results pertaining to the tensor of YIG in Table II. Here we use “VASP” to tag the original tensor7 TABLE II. The force constant tensor of YIG. The columns labeled by A represent the sorted absolute values ofP iKij,αβ and the columns labeled by B list the sorted eigenvalues of Kij,αβ. For the cubic cell of YIG, we obtained the original tensor with the VASP package. Then, we eliminated the el- ements that represent interactions beyond the cutoff radius. This step was done by PASP. Finally, the tensor was modified to meet the requirement of translational symmetry through the optimization formulated in (12). VASP PASP Modified No. A B A B A B 1 0.000 0.000 1.587 -0.102 0.000 0.000 2 0.000 0.000 1.587 -0.102 0.000 0.000 3 0.000 0.000 1.587 -0.102 0.000 0.000 4 0.000 1.065 1.587 0.643 0.000 0.444 5 0.000 1.065 1.587 0.643 0.000 0.444 6 0.000 1.065 1.587 0.643 0.000 0.444 obtained from DFT calculations, “PASP” to label the modified tensor in which interactions beyond the cutoff radius are eliminated, and “Modified” to label the tensor after optimization of independent variables. As shown in Table II, the “PASP” tensor violated the acoustic sum rule and was not positive semi-definite, whereas these is- sues were resolved for the “Modified” tensor. Although an obvious difference existed between the “PASP” and “Modified” tensor in terms of their eigenvalues, we still assumed the target system could be reasonably described by the “Modified” tensor and the validity of this assump- tion would be verified by the calculated results of damp- ing constants. Additional details regarding the selection of tensor elements and the deviation of phonon spectra are provided in Fig. 3. According to figure 3(b) and 3(c), the major deviation in phonon spectra resulted from the elimination of tensor elements, rather than the subse- quent modification. Completing the preparation of Hamiltonian model, we applied the scheme proposed in Sec. II C to our first ob- ject, Y 3Fe5O12. An instance is presented in Figure 4. We setTL= 30K,TS= 180 Kfor the initial nonequilibrium state and adopted an expanded supercell which contains 12800 atoms in the simulation. Fig. 4(a) shows the time evolution of spin temperature in different types of simu- lations. By comparing the curves, we could roughly esti- mate that the equivalent damping constant in SLD simu- lation fell within the range of 10−3∼10−4. To make the estimation more precise, we calculated the initial cool- ing rates ∂TS/∂t|t=0through polynomial (or exponen- tial) fittings and plotted them in Fig. 4(b). Afterwards, a linear regression was performed to determine the quan- titative relation between lg( −∂TS/∂t|t=0) and lg( α). As we expected, the cooling rates in ASD simulations were proportional to the assigned damping constants. Then, we combined the results of SLD and ASD simulations toevaluate the equivalent damping constant. This step was accomplished by identifying the intersection of red and blue lines in Figure 4(b). Finally, the damping constant was determined to be αf= (2.87±0.13)×10−4in this case. To verify our method and result, we present a com- parison between SLD and ASD (where we set α=αf) simulations in Fig. 4(c). The curves agree well with each other in the initial stage but deviate in the second half. This phenomenon is within our expectation, because in the SLD simulation the lattice heats up as the spin cools down, thereby slowing the energy transfer between two subsystems. In addition to the above case, we have measured the equivalent damping constants under different conditions to investigate the temperature dependence of magnetic damping. The final results are summarized in Figure 5. Details about the estimation of uncertainties are given in the supplementary material. For Y 3Fe5O12, the damping constants at different temperatures stay on the order of 10−4, which is in good agreement with the experimental results (3 .2×10−4[47], 2 .2×10−4[48], 1 .2–1.7×10−4 [49]). For example, the damping constant in bulk YIG was reported as 0 .4×10−4in Ref. [50]. Meanwhile, our calculations yielded α= (2.8±0.3)×10−5at ∆T= 15 K and α= (7.0±0.7)×10−5at ∆T= 30 K, where both TL= 0 K. Therefore, the experimental value corresponds roughly to the temperature region of ∆ T= 15∼30 K in our study. We believe such extent of thermal excitation is quite common in all kinds of spintronics experiments. Moreover, Fig. 5 indicates that αis approximately pro- portional to the temperature difference between subsys- tems. This outcome is also consistent with some com- putational works in the past [23, 25]. By comparing the subfigures in Figure 5, we found that αhas little depen- dence on the lattice temperature, although here TLcould be viewed to some extent as the ambient temperature of the spin system. As a supplement to Sec. III A, we further validate our simulations by analyzing the measured cooling rates in Fig. 5(a). By subtracting Eq. (23) from Eq. (22), the transfer rate of energy between magnon and phonon sys- tems can be expressed as, ˙Q=X qpℏωqp⟨˙Nqp⟩=X λ,qpTλ,qp (24) where Tλ,qpdenotes different transfer channels, Tλ,qp∝(nλ−nλ−q)Nqp+nλ−qnλ+ 1 (25) According to the Bose–Einstein distribution, the number of magnons and phonons can be expressed as, nλ=1 eϵλ/kBTS−1, Nqp=1 eℏωqp/kBTL−1(26) When TSis high enough and TLis close to zero, we can approximate nλ=kBTS/ϵλ∝TSandNqpclose to zero. Under these conditions, we have ˙Q∝T2 S. This relation8 FIG. 3. (a) The selection of force constant tensor elements for the cubic cell of YIG. An 160 ×160 zero-one matrix is used to show the result of selection, in which ’1’ denotes the interactions within cutoff radius and ’0’ represents the elements that are artificially eliminated. (b) The phonon spectrum calculated from the force constant tensor before and after the elimination of tensor elements. (c) The phonon spectrum calculated from the force constant tensor before and after the optimization of independent variables. FIG. 4. (a) The time evolution of spin temperature in SLD and ASD simulations. The gray line represents the SLD simulation while the others refer to the ASD simulations with different damping constants. (b) The initial cooling rates ∂TS/∂t|t=0with respect to the damping constants α, where the scaling of axis is set to be logarithm. The gray squares refer to the results of ASD simulations and the blue line acts as the linear regression. The red circle is plotted by intersection of the blue line and the horizontal red dash line, which represents the initial cooling rate in the SLD simulation. Then we can obtain the equivalent damping constant from the abscissa of the red circle. (c) The comparison between ASD and SLD simulations. In the ASD simulation, the Gilbert damping constant is set to be α= 2.87×10−4, which is exactly the result of our evaluation from the SLD simulation. FIG. 5. The temperature dependence of Gilbert damping constants for Y 3Fe5O12. The label of abscissa axis ∆ Trefers to TS−TLof the initial state in dynamical simulations. Measurements on the magnetic damping are performed under different initial conditions of the lattice temperature: (a) TL= 0, (b) TL= 30K, (c)TL= 60K.9 FIG. 6. The relation between damping constants αand spin- lattice coupling constants ∂Jij/∂uk,αin YIG. Through a lin- ear fitting, the slope is determined to be 2 .01, which agrees well with our theoretical predictions. is well verified by linear regressions and the details are provided in the supplementary material. Furthermore, the accuracy of our simulations can also be proved from another perspective. According to Eqs. (22) and (23), the scattering rate Wgrows quadratically with the coupling parameters Mλ,qp. Based on the theory of second quantization, Mλ,qpshall be proportional to the coupling constants ∂Jij/∂uk,α. Therefore, under a definite condition of temperature, we have: α∝˙Q∝∆W∝M2 λ,qp∝(∂Jij/∂uk,α)2(27) In order to verify this relation, we adjusted the spin- lattice coupling constants of YIG coherently while keep- ing the other model parameters unchanged. Then, SLD simulations were carried out to evaluate the correspond- ing damping constants. The result is plotted in Fig. 6, where the x-label “slcc” stands for the spin-lattice cou- pling constants and the subscript “0” refers to the orig- inal situation. From a linear fitting, the slope is deter- mined to be 2 .01, which agrees well with our prediction. C. Damping constants in MnFe 2O4 After the calculation on YIG, we applied our method to MnFe 2O4(MFO), which was reported to possess a large Gilbert damping constant in the literature [13, 51]. As shown in Fig. 7(a), MnFe 2O4has a typical structure of spinels, where A sites are surrounded by four oxygen atoms and B sites are located in octahedrons. Generally, spinels can be classified into normal and inverse struc- tures according to the distribution of divalent and triva- lent cations between A/B sites. In experiments, MFO usually crystallizes into a mixed phase where the normal structure occupies the major part (80% in bulk MFO [52]). Here, we only considered its normal structure in this work. Also, the magnetic ground state of MFO is shown in Fig. 22(b), where the magnetic moments are antiparallel between A/B sites. FIG. 7. (a) The cubic cell of MnFe 2O4. The purple balls rep- resent manganese atoms, the golden balls refer to iron atoms, and the red balls stand for oxygen atoms. (b) The magnetic ground state of MFO. The arrows of different colors repre- sent the spin directions of Mn and Fe atoms separately. (c) The density of states obtained by DFT calculations. (d) The temperature dependence of average magnetization measured in MC and ASD simulations. For MnFe 2O4, the phase tran- sition point from ferrimagnetic to paramagnetic lies in 730K approximately. Firstly, we started to construct an effective Hamilto- nian model for MFO. With the same cutoff settings for YIG, we found 105 nonequivalent interactions, including 4 Heisenberg exchange terms and 10 spin-lattice coupling terms. Subsequently, DFT calculations were carried out to determine the interaction parameters. In these calcu- lations, we adopted a cubic cell containing 56 atoms and a Γ-centered 4 ×4×4 grid mesh in the reciprocal space. Besides, UMn= 3.3 eV and UFe= 3.6 eV were used as the effective Hubbard parameters [52]. With the exception of aforementioned settings, all the relevant first-principles calculations were performed under the same conditions as in Sec. III B. The DOS of MnFe 2O4is plotted in Fig. 7(c), yielding a calculated band gap of 0.612 eV. This value does not match with the result of transport experiments, which re- ported a much smaller band gap (0 .04–0.06 eV) [53]. In addition, MC and ASD simulations were performed using the Heisenberg exchange coefficients listed in Table III. The temperature dependence of average magnetization, shown in Fig. 7(d), suggests the critical temperature to be around 730 K. This result is significantly higher than the measured value of 573 K [54]. Both of the above dis- crepancies may be attributed to the inevitable difference between the ideal normal spinel structure in calculations and the partially disordered samples in reality. Despite this problem, we proceeded to describe the target system with our Hamiltonian model and expected to see how far the calculated results of damping constants would differ10 TABLE III. The exchange coefficients J of MnFe 2O4, where an effective spin S= 1 is adopted. Spin Pair. Distance (Angst) J (meV) 1NN Fe-Fe 3.003 6.835 1NN Mn-Fe 3.521 33.224 1NN Mn-Mn 3.667 3.956 2NN Fe-Fe 5.201 0.929 from experimental values. After the preparation of Hamiltonian model, we con- ducted dynamics simulations to evaluate the equivalent damping parameters in MFO at different temperatures. A supercell containing 13440 atoms was adopted in the simulation, and the results are summarized in Fig. 10. The average of calculated damping constants is around 8×10−5, which is much smaller than the measured value, 1.0×10−2[13, 51]. Two factors may account for this in- consistency. Firstly, the inhomogeneity in real MnFe 2O4 samples greatly enhances the scattering of magnons and phonons, thereby increasing the damping constants. Ad- ditionally, due to the narrow band gap observed in ex- periments, eddy currents can arise at finite temperatures, which leads to a rapid loss of energy in the form of joule heat. As the result of these factors, we failed to obtain a reasonable estimation of Gilbert damping constants for MnFe 2O4with our methodology. On the other side, the contribution of different relaxation mechanisms to FMR linewidth has been studied comprehensively for MnFe 2O4 in Ref. [53], which further confirms our analyses. D. Damping constants in Cr 2O3 Chromia (Cr 2O3) is a well-known collinear magneto- electric antiferromagnet, which holds great prospects in the field of spintronics [55–57]. As shown in Fig. 8(a), the primitive cell of Cr 2O3contains 10 atoms, with each chromium atom bonded to the six oxygen atoms around it. Additionally, Fig. 8(b) displays the magnetic ground state of Cr 2O3, where the spins of two nearest neighbor- ing Cr atoms are oriented in opposite directions. As a preliminary step in constructing the Hamiltonian model, we set the cutoff radius of interactions to be 11.0 Bohr for atomic pairs and 7.0 Bohr for 3-body clusters. Through symmetry analyses, we identified 319 nonequiv- alent interactions, including 5 Heisenberg exchange terms and 21 spin-lattice coupling terms. Afterwards, a series of first-principles calculations were performed to determine the model parameters. Following the settings in Ref. [58], we adopted a hexagonal cell of Cr2O3which contained a total of 90 atoms in the calcula- tions. Additionally, we used the LSDA+U method in its full spherically symmetric form [59]. As to the Hubbard parameters, Jwas fixed at its recommended value of 0.6 FIG. 8. (a) The primitive cell of Cr 2O3. The dark blue balls represent chromium atoms, and the red balls stand for oxygen atoms. (b) The magnetic ground state. The arrows of differ- ent colors represent the spin directions of Cr atoms. (c) The density of states obtained by DFT calculations. (d) The tem- perature dependence of sublattice magnetization measured in MC and ASD simulations. For Cr 2O3, the phase transition point from ferrimagnetic to paramagnetic lies in 310K approx- imately. TABLE IV. The exchange coefficients J of Cr 2O3, in which an effective spin S= 1 is adopted. Spin Pair. Distance (Angst) J (meV) 1NN Cr-Cr 2.640 44.778 2NN Cr-Cr 2.873 29.269 3NN Cr-Cr 3.411 -0.182 4NN Cr-Cr 3.635 0.007 5NN Cr-Cr 4.137 -0.500 eV, and Uwas adjusted to fit the N´ eel temperature ob- served in experiments [60]. We found U= 2.0 eV was the optimal value for 3 delectrons of Cr ions. Except for the settings specified above, all the DFT calculations were conducted under the same conditions as in Sec. III C. The DOS of Cr 2O3is plotted in Fig. 8(c), which yields a calculated band gap of 1.935 eV. This value indicates that the energy dissipation of electric currents can be ne- glected in this system. Additionally, we list the Heisen- berg exchange coefficients of chromia in Table IV. Both MC and ASD simulations were performed to investigate the temperature dependence of sublattice magnetization. According to Fig. 8(d), the critical point was determined to be 310 K approximately, which was quite consistent with experimental observations. Also, the force constants of Cr 2O3went through the modification formulated in Sec. II B, and the spin-lattice coupling parameters are provided in the supplementary material. After the construction of Hamiltonian model, we con- ducted a series of dynamics simulations to evaluate the11 FIG. 9. (a) The 1NN FeT-FeOpair in Y 3Fe5O12. (b) The 1NN Cr-Cr pair in Cr 2O3. The steel blue arrow stands for the orientation of ∂J/∂u and the red number along with it represents the magnitude in unit of meV/Angst. equivalent damping parameters in Cr 2O3. An expanded hexagonal cell containing 14400 atoms was adopted for the simulation, and the results are summarized in Fig. 11. As two specific cases, our calculation yielded α= (1.31± 0.14)×10−4at ∆T= 15 K and α= (2.7±0.3)×10−4 at ∆T= 30 K, where both TL= 0 K. Therefore, the calculated damping constants within ∆ T= 15∼30 K are quite close to 2 ×10−4, which is the estimated value reported in Ref. [61]. Furthermore, the damping constants in Cr 2O3exhibit a significant non-linear relation with the temperature dif- ference of subsystems. Through logarithmic fittings, we calculated the power exponents for Figures 11(a) to 11(c), and the results were 1.17, 1.62, 1.38. If we disregard the difference between ∆ TandTfor the moment, these val- ues are in good agreement with the theoretical prediction of Kasuya and LeCraw [26]. According to their study, the relaxation rate varies as Tnwhere n= 1∼2 while n= 2 corresponds to a larger regime of temperatures. Compared to YIG, the greater magnetic damping ob- served in chromia can be attributed to its significantly stronger spin-lattice coupling. As shown in Fig. 9, the magnitude of principal spin-lattice coupling constant in Cr2O3is two or three times larger than that in YIG. This could be explained by the fact that direct exchange in- teraction between two magnetic atoms decreases rapidlywith their distance [62]. Therefore, owing to the shorter distance of Cr-Cr pair, the direct exchange interaction between neighboring Cr atoms is believed to have a great contribution to the spin-lattice coupling in Cr 2O3. IV. CONCLUSIONS In summary, we propose a scheme to evaluate the con- tribution of spin-lattice coupling to the Gilbert damp- ing in insulating magnetic materials. Our methodology involves first-principles based Hamiltonian models and spin-lattice dynamics simulations. Following a series of validations, we applied our method to three magnetic ma- terials, namely Y 3Fe5O12, MnFe 2O4and Cr 2O3. Their damping constants were estimated separately, and the results show that, in general, αis approximately propor- tional to the temperature difference between spin and lattice subsystems. Under the condition of ∆ T= 30 K, the calculated damping constants are averaged to be 0.8×10−4for YIG, 0 .2×10−4for MFO and 2 .2×10−4 for Cr 2O3. The results for YIG and Cr 2O3are in good agreement with experimental measurements, while the discrepancy for MFO can be attributed to the inhomo- geneity and small band gap in real samples. Overall, the approach presented in this work holds great promise for accurately predicting the Gilbert damping constants for magnetic insulators. ACKNOWLEDGMENTS This work is supported by the National Key R&D Program of China (No. 2022YFA1402901 ), the Na- tional Natural Science Foundation of China (Grant Nos. 11825403, 11991061, and 12188101), the Guangdong Ma- jor Project of the Basic and Applied Basic Research (Future functional materials under extreme conditions– 2021B0301030005). [1] S. Wolf, D. Awschalom, R. Buhrman, J. Daughton, S. von Molnar, M. Roukes, A. Chtchelkanova, and D. Treger, Spintronics: A spin-based electronics vision for the future, SCIENCE 294, 1488 (2001). [2] T. Jungwirth, X. Marti, P. Wadley, and J. Wunderlich, Antiferromagnetic spintronics, NATURE NANOTECH- NOLOGY 11, 231 (2016). [3] L. Smejkal, Y. Mokrousov, B. Yan, and A. H. MacDon- ald, Topological antiferromagnetic spintronics, NATURE PHYSICS 14, 242 (2018). [4] I. Zutic, J. Fabian, and S. Das Sarma, Spintronics: Fun- damentals and applications, REVIEWS OF MODERN PHYSICS 76, 323 (2004). [5] V. Cherepanov, I. Kolokolov, and V. L’vov, The saga of yig: Spectra, thermodynamics, interaction and relax-ation of magnons in a complex magnet, Physics Reports 229, 81 (1993). [6] A. A. Serga, A. V. Chumak, and B. Hillebrands, Yig magnonics, Journal of Physics D: Applied Physics 43, 264002 (2010). [7] H. Wang, C. Du, P. C. Hammel, and F. Yang, Spin trans- port in antiferromagnetic insulators mediated by mag- netic correlations, Phys. Rev. B 91, 220410 (2015). [8] L. D. Landau and E. M. Lifshitz, On the theory of the dis- persion of magnetic permeability in ferromagnetic bodies, Physikalische Zeitschrift der Sowjetunion 8, 153 (1935). [9] T. Gilbert, A phenomenological theory of damping in fer- romagnetic materials, IEEE Transactions on Magnetics 40, 3443 (2004).12 FIG. 10. The temperature dependence of Gilbert damping constants for MnFe 2O4. The label of abscissa axis ∆ Trefers to TS−TLof the initial state in dynamical simulations. Measurements on the magnetic damping are performed under different initial conditions of the lattice temperature: (a) TL= 0, (b) TL= 30K, (c)TL= 60K. FIG. 11. The temperature dependence of Gilbert damping constants for Cr 2O3. The label of abscissa axis ∆ Trefers to TS−TL of the initial state in dynamical simulations. Measurements on the magnetic damping are performed under different initial conditions of the lattice temperature: (a) TL= 0, (b) TL= 30K, (c)TL= 60K. [10] S. Karakurt, R. Chantrell, and U. Nowak, A model of damping due to spin–lattice interaction, Journal of Mag- netism and Magnetic Materials 316, e280 (2007), pro- ceedings of the Joint European Magnetic Symposia. [11] M. C. Hickey and J. S. Moodera, Origin of intrinsic gilbert damping, Phys. Rev. Lett. 102, 137601 (2009). [12] A. Widom, C. Vittoria, and S. Yoon, Gilbert ferromag- netic damping theory and the fluctuation-dissipation the- orem, Journal of Applied Physics 108, 073924 (2010). [13] C. Vittoria, S. D. Yoon, and A. Widom, Relaxation mech- anism for ordered magnetic materials, Phys. Rev. B 81, 014412 (2010). [14] O. Eriksson, A. Bergman, L. Bergqvist, and J. Hellsvik, Atomistic spin dynamics: Foundations and applications (Oxford university press, 2017). [15] R. F. L. Evans, W. J. Fan, P. Chureemart, T. A. Ostler, M. O. A. Ellis, and R. W. Chantrell, Atomistic spin model simulations of magnetic nanomaterials, Journal of Physics: Condensed Matter 26, 103202 (2014). [16] V. Kambersk´ y, On ferromagnetic resonance damping in metals, Czechoslovak Journal of Physics B 26, 1366 (1976). [17] D. Steiauf and M. F¨ ahnle, Damping of spin dynamics in nanostructures: An ab initio study, Phys. Rev. B 72, 064450 (2005).[18] K. Gilmore, Y. U. Idzerda, and M. D. Stiles, Identifica- tion of the dominant precession-damping mechanism in fe, co, and ni by first-principles calculations, Phys. Rev. Lett. 99, 027204 (2007). [19] V. Kambersk´ y, Spin-orbital gilbert damping in common magnetic metals, Phys. Rev. B 76, 134416 (2007). [20] A. Brataas, Y. Tserkovnyak, and G. E. W. Bauer, Scat- tering theory of gilbert damping, Phys. Rev. Lett. 101, 037207 (2008). [21] H. Ebert, S. Mankovsky, D. K¨ odderitzsch, and P. J. Kelly, Ab initio calculation of the gilbert damping parameter via the linear response formalism, Physical Review Let- ters107, 066603 (2011). [22] S. Mankovsky, D. K¨ odderitzsch, G. Woltersdorf, and H. Ebert, First-principles calculation of the gilbert damp- ing parameter via the linear response formalism with ap- plication to magnetic transition metals and alloys, Phys- ical Review B 87, 014430 (2013). [23] M. Aßmann and U. Nowak, Spin-lattice relaxation be- yond gilbert damping, Journal of Magnetism and Mag- netic Materials 469, 217 (2019). [24] J. Tranchida, S. Plimpton, P. Thibaudeau, and A. Thompson, Massively parallel symplectic algorithm for coupled magnetic spin dynamics and molecular dy- namics, Journal of Computational Physics 372, 40613 (2018). [25] M. Strungaru, M. O. A. Ellis, S. Ruta, O. Chubykalo- Fesenko, R. F. L. Evans, and R. W. Chantrell, Spin- lattice dynamics model with angular momentum transfer for canonical and microcanonical ensembles, Phys. Rev. B103, 024429 (2021). [26] T. Kasuya and R. C. LeCraw, Relaxation mechanisms in ferromagnetic resonance, Phys. Rev. Lett. 6, 223 (1961). [27] F. Lou, X. Li, J. Ji, H. Yu, J. Feng, X. Gong, and H. Xi- ang, Pasp: Property analysis and simulation package for materials, The Journal of Chemical Physics 154, 114103 (2021). [28] H. J. Xiang, E. J. Kan, S.-H. Wei, M.-H. Whangbo, and X. G. Gong, Predicting the spin-lattice order of frus- trated systems from first principles, Phys. Rev. B 84, 224429 (2011). [29] H. Xiang, C. Lee, H.-J. Koo, X. Gong, and M.- H. Whangbo, Magnetic properties and energy-mapping analysis, Dalton Transactions 42, 823 (2013). [30] E. Prodan and W. Kohn, Nearsightedness of electronic matter, Proceedings of the National Academy of Sciences 102, 11635 (2005). [31] T. W. Ko, J. A. Finkler, S. Goedecker, and J. Behler, A fourth-generation high-dimensional neural network po- tential with accurate electrostatics including non-local charge transfer, Nature communications 12, 1 (2021). [32] A. Togo and I. Tanaka, First principles phonon calcu- lations in materials science, Scripta Materialia 108, 1 (2015). [33] G. Kresse and J. Hafner, Ab initio molecular dynamics for open-shell transition metals, Phys. Rev. B 48, 13115 (1993). [34] G. Kresse and J. Furthm¨ uller, Efficient iterative schemes for ab initio total-energy calculations using a plane-wave basis set, Phys. Rev. B 54, 11169 (1996). [35] G. Kresse and J. Furthm¨ uller, Efficiency of ab-initio total energy calculations for metals and semiconductors using a plane-wave basis set, Computational Materials Science 6, 15 (1996). [36] P. Virtanen, R. Gommers, T. E. Oliphant, M. Haber- land, T. Reddy, D. Cournapeau, E. Burovski, P. Peter- son, W. Weckesser, and J. Bright, Scipy 1.0: fundamen- tal algorithms for scientific computing in python, Nature methods 17, 261 (2020). [37] W. B. Nurdin and K.-D. Schotte, Dynamical temperature for spin systems, Phys. Rev. E 61, 3579 (2000). [38] J. H. Mentink, M. V. Tretyakov, A. Fasolino, M. I. Kat- snelson, and T. Rasing, Stable and fast semi-implicit in- tegration of the stochastic landau–lifshitz equation, Jour- nal of Physics: Condensed Matter 22, 176001 (2010). [39] N. Grønbech-Jensen and O. Farago, A simple and effec- tive verlet-type algorithm for simulating langevin dynam- ics, Molecular Physics 111, 983 (2013). [40] D. Wang, J. Weerasinghe, and L. Bellaiche, Atomistic molecular dynamic simulations of multiferroics, Phys. Rev. Lett. 109, 067203 (2012). [41] K. P. Sinha and U. N. Upadhyaya, Phonon-magnon inter- action in magnetic crystals, Phys. Rev. 127, 432 (1962). [42] U. N. Upadhyaya and K. P. Sinha, Phonon-magnon inter- action in magnetic crystals. ii. antiferromagnetic systems, Phys. Rev. 130, 939 (1963). [43] D. Campbell, C. Xu, T. Bayaraa, and L. Bellaiche, Finite-temperature properties of rare-earth iron garnets in a magnetic field, Phys. Rev. B 102, 144406 (2020).[44] P. E. Bl¨ ochl, Projector augmented-wave method, Phys. Rev. B 50, 17953 (1994). [45] J. P. Perdew, A. Ruzsinszky, G. I. Csonka, O. A. Vydrov, G. E. Scuseria, L. A. Constantin, X. Zhou, and K. Burke, Restoring the density-gradient expansion for exchange in solids and surfaces, Phys. Rev. Lett. 100, 136406 (2008). [46] S. L. Dudarev, G. A. Botton, S. Y. Savrasov, C. J. Humphreys, and A. P. Sutton, Electron-energy-loss spec- tra and the structural stability of nickel oxide: An lsda+u study, Phys. Rev. B 57, 1505 (1998). [47] Y. Sun, Y.-Y. Song, H. Chang, M. Kabatek, M. Jantz, W. Schneider, M. Wu, H. Schultheiss, and A. Hoff- mann, Growth and ferromagnetic resonance properties of nanometer-thick yttrium iron garnet films, Applied Physics Letters 101, 152405 (2012). [48] M. Onbasli, A. Kehlberger, D. H. Kim, G. Jakob, M. Kl¨ aui, A. V. Chumak, B. Hillebrands, and C. A. Ross, Pulsed laser deposition of epitaxial yttrium iron garnet films with low gilbert damping and bulk-like magnetiza- tion, APL Materials 2, 106102 (2014). [49] C. Dubs, O. Surzhenko, R. Linke, A. Danilewsky, U. Br¨ uckner, and J. Dellith, Sub-micrometer yttrium iron garnet lpe films with low ferromagnetic resonance losses, Journal of Physics D: Applied Physics 50, 204005 (2017). [50] P. R¨ oschmann and W. Tolksdorf, Epitaxial growth and annealing control of fmr properties of thick homogeneous ga substituted yttrium iron garnet films, Materials Re- search Bulletin 18, 449 (1983). [51] J. Goodenough, W. Gr¨ aper, F. Holtzberg, D. Huber, R. Lefever, J. Longo, T. McGuire, and S. Methfessel, Magnetic and other properties of oxides and related com- pounds (Springer, 1970). [52] J.-R. Huang and C. Cheng, Cation and magnetic orders in mnfe2o4 from density functional calculations, Journal of Applied Physics 113, 033912 (2013). [53] A. G. Flores, V. Raposo, L. Torres, and J. I˜ niguez, Two- magnon processes and ferrimagnetic linewidth calcula- tion in manganese ferrite, Phys. Rev. B 59, 9447 (1999). [54] P. V. Reddy, K. Pratap, and T. S. Rao, Electrical con- ductivity of some mixed ferrites at curie point, Crystal Research and Technology 22, 977 (1987). [55] N. Hedrich, K. Wagner, O. V. Pylypovskyi, B. J. Shields, T. Kosub, D. D. Sheka, D. Makarov, and P. Maletinsky, Nanoscale mechanics of antiferromagnetic domain walls, Nature Physics 17, 574 (2021). [56] W.-Z. Xiao, Y.-W. Zhang, L.-L. Wang, and C.-P. Cheng, Newtype two-dimensional cr2o3 monolayer with half-metallicity, high curie temperature, and magnetic anisotropy, Journal of Magnetism and Magnetic Mate- rials543, 168657 (2022). [57] P. Makushko, T. Kosub, O. V. Pylypovskyi, N. Hedrich, J. Li, A. Pashkin, S. Avdoshenko, R. H¨ ubner, F. Ganss, D. Wolf, et al. , Flexomagnetism and vertically graded n´ eel temperature of antiferromagnetic cr2o3 thin films, Nature Communications 13, 6745 (2022). [58] S. Shi, A. L. Wysocki, and K. D. Belashchenko, Mag- netism of chromia from first-principles calculations, Phys. Rev. B 79, 104404 (2009). [59] A. I. Liechtenstein, V. I. Anisimov, and J. Zaanen, Density-functional theory and strong interactions: Or- bital ordering in mott-hubbard insulators, Phys. Rev. B 52, R5467 (1995). [60] C. A. Brown, Magneto-electric domains in single crystal chromium oxide , Ph.D. thesis, Imperial College London14 (1969). [61] K. D. Belashchenko, O. Tchernyshyov, A. A. Kovalev, and O. A. Tretiakov, Magnetoelectric domain wall dy- namics and its implications for magnetoelectric memory,Applied Physics Letters 108, 132403 (2016). [62] J. Li, J. Feng, P. Wang, E. Kan, and H. Xiang, Na- ture of spin-lattice coupling in two-dimensional cri3 and crgete3, SCIENCE CHINA-PHYSICS MECHANICS & ASTRONOMY 64, 10.1007/s11433-021-1717-9 (2021). | 2023-09-20 | Magnetic damping has a significant impact on the performance of various
magnetic and spintronic devices, making it a long-standing focus of research.
The strength of magnetic damping is usually quantified by the Gilbert damping
constant in the Landau-Lifshitz-Gilbert equation. Here we propose a
first-principles based approach to evaluate the Gilbert damping constant
contributed by spin-lattice coupling in magnetic insulators. The approach
involves effective Hamiltonian models and spin-lattice dynamics simulations. As
a case study, we applied our method to Y$_3$Fe$_5$O$_{12}$, MnFe$_2$O$_4$ and
Cr$_2$O$_3$. Their damping constants were calculated to be $0.8\times10^{-4}$,
$0.2\times10^{-4}$, $2.2\times 10^{-4}$, respectively at a low temperature. The
results for Y$_3$Fe$_5$O$_{12}$ and Cr$_2$O$_3$ are in good agreement with
experimental measurements, while the discrepancy in MnFe$_2$O$_4$ can be
attributed to the inhomogeneity and small band gap in real samples. The
stronger damping observed in Cr$_2$O$_3$, compared to Y$_3$Fe$_5$O$_{12}$,
essentially results from its stronger spin-lattice coupling. In addition, we
confirmed a proportional relationship between damping constants and the
temperature difference of subsystems, which had been reported in previous
studies. These successful applications suggest that our approach serves as a
promising candidate for estimating the Gilbert damping constant in magnetic
insulators. | Evaluating Gilbert Damping in Magnetic Insulators from First Principles | 2309.11152v1 |
Enhanc ed Gilbert Damping in Re doped FeCo Films – a combined experimental and theoretical study S. Akansel1, A. Kumar1, V.A.Venugopal2, R.Banerjee3, C. Autieri3, R.Brucas1, N. Behera1, M. A. Sortica3, D. Primetzhofer3, S. Basu2, M.A. Gubbins2, B. Sanyal3, and P. Svedlindh1 1Department of Engineering Sciences , Uppsala University, Box 534, SE -751 21 Uppsala, Sweden 2Seagate Technology, BT48 0BF, Londonderry, United Kingdom 3Department of Physics and Astronomy, Uppsala University, Box 516, SE -751 20 Uppsala, Sweden The effect s of rhenium doping in the range 0 – 10 at% on the static and dynamic magnetic properties of Fe65Co35 thin films have been studied experimentally as well as with first principles electronic structure calculations focussing on the change of the saturation magnetization (𝑀𝑠) and the Gilbert damping parameter ( 𝛼) Both experiment al and theoretical results show that 𝑀𝑠 decreases with increasing Re doping level, while at the same time 𝛼 increases. The experimental low temperature saturation magnetic induction exhibits a 2 9% decrease, from 2.3 1T to 1. 64T, in the investigated doping concentration range , which is more than predicted by the theoretical calculations. The room temperature value of the damping parameter obtained from ferromagnetic resonance measurements , correcting for extrinsic contributions to the damping, is for the undoped sample 2.7×10−3, which is close to the theoretically calculated Gilbert damping parameter . With 10 at% Re doping , the damping parameter increases to 9.0×10−3, which is in good agreement with the theoretical value of 7.3×10−3. The increase in damping parameter with Re doping is explained by the increase in density of states at Fermi level, mostly contributed by the s pin-up channel of Re. Moreover, both experimental and theoretical values for the da mping parameter are observed to be weakly decreas ing with decreasing temperature . 1. INTRODUCTION During the last decades , thin films of soft magnetic alloys such as NiFe and FeCo have been in focus due to possible use in applications such as spin valves ,1,2 magnetic tunneling junctions ,3,4,5 spin injectors ,6 magnetic storage technologies and in particular in magnetic recording write heads .7 Beside s spintronic and magnetic memory devices , such materials are useful for shielding applications that are necessary in order to reduce the effect of electromagnetic fields created by electronic devices. The magnetic damping parameter of the material play s a critical role for the performance of such spintronic and memory devices as well as for shielding applications. On the one hand, a low damping parameter is desired in order to get low critical switching current in spintronic devices .8,9,10 On the other hand , a high damping parameter is necessary in order to reduce the magetization switching time in magnetic memory devices and to be able to operate devices at high speeds .11 FeCo alloys are promising materials for high frequency spintronic applications and magnetic recording devices due to their high saturation magnetization (𝑀𝑠), high permeability, thermal stability and comparably high resistivity .12,13,14 One possible drawback is that FeCo alloy s exhibit high coercivity (𝐻𝑐), which is not favorable for such applications , however this problem can be solved by thin film growth on suitable buffer layer s.15,16,12 Except coercivity problems, the damping parameter of these materials should be increased to make them com patible for high speed devices . Dynamic properties of magnetic materials are highly dependent on the damping parameter. This parameter is composed of both intrinsic and extrinsic contributions. The intrinsic contribution is called the Gilbert damping and depends primarily on the spin-orbit coupling .17 Intrinsic damping is explained as scattering of electrons by phonons and magnons .18,19 Beside s electron scattering , due to the close relation between magnetocrystalline anisotropy and spin-orbit coupling , it can be assumed that the intrinsic damping is also related to the magnetocrystalline anisotropy constant .20 Regarding extrinsic damping , there can be a number of different contributions. The most common contribution originates from two magnon scattering (TMS) .21 However , this contribution vanishes when ferromagnetic resonance (FMR) measurements are performed by applying the static magnetic field along the film normal in inplane anisotropic thin films .22 Beside s TMS , there are some other extrinsic contributions to the damping that are not possible to get rid of by changing the measurement configuration . One of these contributions is radiative damping , which arises from inductive coupling between the precessing magnetization and the waveguide used for FMR measurem ents.23 Another contribution for metallic ferromagnetic films is the eddy current damping related to microwave magnetic field induced eddy currents in the thin film s during measurement s.23,24 In order to make a soft magnetic thin film suitible for a specific applica tion, taking into account requirements set by the device application , its damping paramete r should be tailored. As mentioned above , an increased damping parameter is necesssary for devices requiring high switching speed . Several efforts have been made for enhanching the damping parameter of soft magnetic materials. NiFe alloys constitu te one of the most studied systems in this respect . The most common way to enhance the intrinsic damping of an all oy is to dope it with differ ent elements . Rare earth elements with large spin-orbit coupling have revealed promising results as dopant s in terms of increas ed dampin g parameter .25,26,27 3d, 4d and 5d transition metals dopants have also been studied experimentally , revealing an increase of the damping parameter .28,29 Beside s experimental results , theoretical calculations support the idea that transition metals and especially 5d elements can enhance the damping parameter of NiF e alloys due to scattering in presence of chemical disorde r , as well as due to the effect of spin -orbit coupling .30 Although NiFe alloys have been the focus in several extensive studies, FeCo alloys have so far not been studied to the same extent . Attempts have been made to dope FeCo with Yb,20 Dy,31 Gd,32 and Si ,33 where in all cases an increase of the damping parameter was observed . Apart from doping of alloys , the addition of adjacent layers to NiFe and CoFe has also been studied . In particular , adding layers consisting of rare earth elem ents with large orbital moment s gave positive results in terms of increased damping parameter .34 Fe65Co35 alloy s are attractive material s because of high 𝑀𝑠 and reduced 𝐻𝑐 values. However , not much is known about the magnetic damping mechanism s for this composition . Since it is of interest for high data rate magnetic memory devices, the damping parameter should be increased in order to make the magnetic switching faster. To the best of our knowledge , systematic doping of Fe 65Co35 with 5d elements has not been studied so far experimentally . Some of us have found from ab initio calculations that 5d transition metal dopants can increase the damping parameter and Re is one of the potential candidates.35 Re is particularly interesting as it has a nice compromise of having not so much reduced saturation magnetization and a quite enh anced damping parameter. In this work, we have perfomed a systematic ab initio study of Fe65Co35 doped with increasing Re concentration to find an increasing damping parameter . The theoretical prediction s are confirmed by results obtained from temperature dependent FMR measurements performed on Re doped Fe65Co35 films. 2. EXPERIMENTAL AND THEORETIC AL METHOD S Rhenium doped Fe 65Co35 samples were prepared by varying the Re concentration from 0 to 10.23 at%. All samples were deposited using DC magnetron sputtering on Si/SiO 2 substrate s. First a 3 nm thick Ru seed layer was deposited on the Si/SiO 2 substrate followed by room temperature deposition of 20 nm and 40 nm thick Re -doped Fe65Co35 films by co -sputtering between Fe 65Co35 and Re target s. Finally, a 3 nm thick Ru layer was deposited as a capping layer over the Re -doped Fe65Co35 film. The nominal Re concentration was derived from the calibrated deposition rate used in the deposition system. The nominal Re doping concentration s of the Fe65Co35 samples are as follows ; 0, 2.62, 5.45 and 10.23 at%. The crystalline structure of the fims were investigated by utilizing grazing incident X -Ray diffraction (GIXRD). The i ncidence angle was fixed at 1o during GIXRD measurements and a CuKα source was used. Accurate values for film thickness and interface roughness were determined by X -ray reflectivity (XRR) measurements. Beside XRD , composition and areal density of the films were deduced by Rutherford backscattering spectrometry36 (RBS) with ion beams of 2 MeV 4He+ and 10 MeV 12C+. The beams were provided by a 5 MV 15SDH -2 tandem accelerator at the Tandem Laboratory at Uppsala University. The experiments were performed with the incident beam at 5° with respect to the surface normal and scattering angles of 170° and 120° . The experimental data was evaluated with the SIMNRA program .37 In-plane magnetic hysteresis measurments were performed using a Magnetic Property Measurement System (MPMS, Quantum Design) . Ferromagnetic resonance measurements were performed using two different techniques. First in- plane X -band (9.8 GHz) cavity FMR measurements were performed . The setup is equipped with a goniometer making it possible to rotate the sample with respect to the applied magnetic field; in this way the in -plane anisotropy fields of the different samples have been determine d. Beside s cavity FMR studies , a setup for broadband out-of-plane FMR measurements have been utilized . For out -of-plane measurements a vector network analyzer (VNA) was used. Two ports of the VNA were connected to a coplanar waveguide (CPW) mounted on a Ph ysical Property Measurement System (PPMS, Quantum Design) multi -function probe . The PPMS is equipped with a 9 T superconducting magnet, which is needed to saturate Fe65Co35 films out -plane and to detect the FMR signal. The broadband FMR measurements were carried out a t a fixed microwave frequency using the field -swept mode, repeating the measurement for different f requencies in the range 15 – 30GHz. The theoretical calculations are based on spin -polarized relativistic m ultiple scattering theory using the Korringa -Kohn -Rostoker (KKR) formalism implemented in the spin polarized relativistic KKR code (SPR-KKR) . The Perdew -Burke -Ernzerhof (PBE) exchange -correlation functional within generalized gradient approximation was used. The equilibrium lattice parameter s were obtained by energy minimization for each composition. Substitutional disorder was treated within the Coherent Potential Approximation (CPA). The damping parameters were calcu lated by the method proposed by Mankovsky et al.,38 based on the ab initio Green's function technique and linear res ponse formalism where one takes into consi deration scattering processes as well as spin - orbit coupling built in Dirac's relativistic formulation. The calculations of Gilbert damping parameters at finite temperatures were done using an alloy -analo gy model of atomic displacements corresponding to the thermal average of the root mean square displacement at a given temperature. 3. RESULTS AND DISCUSSION Re concentrations and layer thickness (areal densities) of the 20 nm doped films were obtained by RBS experiments. RBS employing a beam of 2 MeV He primary ions was used to deduce the areal concentration of each layer. Additional measurements with 10 MeV C probing particles permit to resolve the atomic fractions of Fe, Co and Re. The spectra for the samples with different Re concentration are shown in Fig. A1 . The measured Re concentrations are 3.0±0.1 at%, 6.6±0.3 at% and 12.6±0.5 at%. Moreover, the results for Fe and Co atomic fractions show that there is no preferential replacement by Re , implying that the two elements are replaced according to their respective concentration . Figure 1 (a) shows GIXRD spectra in the 2𝜃-range from 20o to 120o for the Fe65Co35 films with different Re concentration. Diffraction peaks corresponding to the body centered cubic Fe 65Co35 structure have been indexed in the figure; no other diffraction peaks appear in the different spectra. Depending on the Re -dopant concentration shi fts in the peak positions are observed, the diffraction peaks are suppressed to lower 2𝜃-values with increasing dopant concentration . The shift for the (110) peak for the different dopant concentrations is given as an inset in Fig. 1 (a). Similar shifts are observed for the other diffraction peaks. This trend in peak shift is an experimental evidence of an increasing amount of Re dopant within the deposited thin films. Since the peaks are shifted towards lower 2𝜃-values with increasing amount of Re dopant , the lattice parameter increases with increasing Re concentration.39 Figure 1 (b) shows the experimental as well as theoretically calculated lattice parameter versus Re concentration. The qualitative agreement between theory and experiment is obtained. However, t he rate of lattice parameter increase with increasing Re concentration is larger for the theoretically calculated lattice parameter. This is not unexpected as the generalized gradie nt approximation for the exchange -correlation potential has a tendency to overestimate the lattice parameter. Another possible explanation for the difference in lattice parameter is that the increase of the lattice parameter for the Re -doped Fe 65Co35 films is held back by the compressive strain due to lattice mismatch with Si/SiO 2/Ru. XRR measurements revealed that the surface roughness of the Fe 65Co35 films is less than 1 nm , which cannot affect static and magnetic properties drastically. Results from XRR measurements are given in table 1. Room temperature normalized magnetization curves for the Re-doped Fe 65Co35 films are shown in Fig. 2 (a) . The coercivity for all films is in the range of 2 mT and all films, except for the 1 2.6 at% Re doped film that show a slightly rounded hysteresis loop, exhibit rectangular hysteresis loops. The low value for the coercivity is expected for seed layer grown films .15 The experimentally determined low temperature saturation magnetization together with the theoretically calculated magnetization versus Re concentra tion are shown in Fig. 2 (b). As expected, both experimental and theoretical r esults show that the saturation magnetization decreases with increasing Re concentration . A linear decrease in magnetization is observed in the theoretical calculations whereas a non -linear behavior is seen in the experimental data. Angle resolved cavity FMR measurements were used to study the in -plane magnetic anisotropy . The angular -dependent resonant field ( 𝐻𝑟) data was analyzed using the following equation ,40 𝑓=µ0𝛾 2𝜋[{𝐻𝑟cos(𝜙𝐻−𝜙𝑀)+𝐻𝑐 2cos4(𝜙𝑀−𝜙𝐶)+𝐻𝑢cos2(𝜙𝑀−𝜙𝑢)}{𝐻𝑟cos(𝜙𝐻− 𝜙𝑀)+𝑀𝑒𝑓𝑓+𝐻𝑐 8(3+cos4(𝜙𝑀−𝜙𝐶))+𝐻𝑢cos2(𝜙𝑀−𝜙𝑢)}]12⁄ , (1) where 𝑓 is the cavity resonance frequency and 𝛾 is the gyromagnetic ratio . 𝜙𝐻, 𝜙𝑀, 𝜙𝑢 and 𝜙𝐶 are the in -plane directions for the magnetic field, magnetization, uniaxial anisotropy and cubic anisotropy, respectively, with respect to the [100 ] direction of the Si substrate. 𝐻𝑢=2𝐾𝑢 µ0𝑀𝑠 and 𝐻𝑐=4𝐾𝑐 µ0𝑀𝑠 are the uniaxial and cubic anisotropy fields, where 𝐾𝑢 and 𝐾𝑐 are the uniaxial and cubic magnetic anisotropy constants , and 𝑀𝑒𝑓𝑓 is the effective magnetization. Fitting parameters were limited to 𝑀𝑒𝑓𝑓, 𝛾 and 𝐻𝑢, since the Hr versus ϕH curves did not give any indication of a cubic anisotropy. Figure 3 shows 𝐻𝑟 versus 𝜙𝐻 extracted from the angular -dependent FMR measurements together with fits according to Eq. (1), clearly revealing dominant twofold uniaxial in -plane magnetic anisotropy. Extracted anisotropy field and effective magnetization values are given in Table 2 . The results show that 𝐻𝑢 is within the accuracy of the experiment independent of the Re concentration . Temperature dependent o ut-of-plane FMR measurements were performed in the temperature range 50 K to 300 K recording the complex transmission coefficient 𝑆21. Typical field -swept results for the r eal and imaginary components of 𝑆21 for the undoped and 1 2.6 at% Re-doped samples are shown in Fig. 4. The field -dependent 𝑆21 data was fitted to the following set of equations,41 𝑆21(𝐻,𝑡)=𝑆210+𝐷𝑡+𝜒(𝐻) 𝜒̃0 𝜒(𝐻)=𝑀𝑒𝑓𝑓(𝐻−𝑀𝑒𝑓𝑓) (𝐻−𝑀𝑒𝑓𝑓)2−𝐻𝑒𝑓𝑓2−𝑖𝛥𝐻 (𝐻−𝑀𝑒𝑓𝑓) . (2) In these equations 𝑆210 corresponds to the non-magnetic contribution to the complex transmission signal , 𝜒̃0 is an imaginary function of the microwave frequency and film thickness and 𝜒(𝐻) is the complex susceptibility of the magnetic film. The term 𝐷𝑡 accounts for a linear drift of the recorded 𝑆21 signal. 𝑀𝑒𝑓𝑓=𝑀𝑠−𝐻𝑘⫠, where 𝐻𝑘⫠ is the perpendicular anisotropy field and 𝐻𝑒𝑓𝑓=2𝑓 𝛾µ0. The 𝑆21 spectra were fitted to Eq. (2 ) in order to extract the linewidth 𝛥𝐻 and 𝐻𝑟 values. Fits t o Eq. (2) are shown as solid lines in Fig. 4. The experimentally measured total d amping parameter ( 𝛼𝑡𝑜𝑡𝑎𝑙 ), including both the intrinsic contribution (Gilbert damping) and extrinsic contributions , was extracted by fitting 𝛥𝐻 versus frequency to the following expression, 41 µ0𝛥𝐻=4𝛼𝑡𝑜𝑡𝑎𝑙 𝑓 𝛾+µ0𝛥𝐻0 , (3) where 𝛥𝐻0 is the frequency independent linewidth broadening due to sample inhomogeneity . Beside s 𝛼𝑡𝑜𝑡𝑎𝑙 , 𝑀𝑒𝑓𝑓 can also be extracted by fitting the 𝐻𝑟 versus frequency results to the expression µ0𝐻𝑟=2𝜋𝑓 𝛾+µ0𝑀𝑒𝑓𝑓 . (4) Typical temperature dependent results for 𝑓 versus 𝐻𝑟 and 𝛥𝐻 versus 𝑓 are shown in Fig. 5 for the 1 2.6 at% Re -doped Fe65Co35 film. Extracted values of 𝑀𝑒𝑓𝑓 at different temperatures are given in Table 3 for all samples . As expected, the results show that 𝑀𝑒𝑓𝑓 decreas es with increasing dopant concentration. Since 𝑀𝑒𝑓𝑓=𝑀𝑠−𝐻𝑘⫠ and the film thickness is large enough to make a possible contribution from out -of-plane anisotropy negligible one can make the justified assumption that 𝑀𝑒𝑓𝑓≈𝑀𝑠. The analysis using Eqs. (2) – (4) also give values for the Land é 𝑔- factor ( 𝛾=𝑔µ𝐵 ħ), yielding 2.064 and 2.075 for the undoped and 12.6 at% doped samples, respectively (similar values are obtained at all temperatures). As indicated above, the d amping parameters extracte d from FMR measurements ( 𝛼𝑡𝑜𝑡𝑎𝑙) include both intrinsic and extrinsic contributions. One of the most common extrinsic contribution s is TMS , which is avoided in this study by measuring FMR with the magnetic field applied out of the film plane. Except TMS , extrinsic contributions such as eddy curr ent damping and radiative damping are expected to contribute the measured damping . In a metallic ferromagnet, which is placed on top of a CPW , precession of spin waves induces AC currents in the ferromagnet ic film, thereby dissipating energy . The radiative damping has similar origin as the eddy current damping, but here the precession of the magnetization induces microwave -frequency currents in the CPW where energy is dissipated. Thus, there are two extrinsic contributions to the measured damping ; one that is caused by eddy currents in the ferromagnet ic film (𝛼𝑒𝑑𝑑𝑦) and another one caused by eddy currents in the CPW ( 𝛼𝑟𝑎𝑑).23 In order to obtain the reduced damping of the films (𝛼𝑟𝑒𝑑), which we expect to be close to the intrinsic damping of the films, the extrinsic contributions should be subtracted from 𝛼𝑡𝑜𝑡. We have neglected any contribution to the measured damping originating from spin -pumping into seed and capping layers. However, since spin -pumping in low spin -orbit coupling materials like Ru with thickness quite less than the spin -diffusion length is quit e small, the assumption of negligible contribution from spin -pumping is justified. The t otal damping can thus be given as 𝛼𝑡𝑜𝑡=𝛼𝑟𝑒𝑑+𝛼𝑟𝑎𝑑+𝛼𝑒𝑑𝑑𝑦 . When the precession of the magnetization is assumed to be uniform in the sample , the expression for radiative damping can be given as23 𝛼𝑟𝑎𝑑=𝜂𝛾µ02𝑀𝑠𝛿𝑙 2𝑍0𝑤 , (5) where 𝑍0 =50 Ω is the waveguide impedance, 𝑤=240 µm is the width of the CPW center conductor , 𝜂 is a dimensionless parameter that accounts for FMR mode profile, δ is the thickness and 𝑙 is the length of the sample. The l ength of all samples were 4mm and the thickness 20nm for the undoped and 1 2.6 at% Re-doped films and 40nm for the 3.0 at% and 6.6 at% Re-doped films. Temperature dependent radiative damping contributions for all Fe 65Co35 films are given in Table 4. Beside s 𝛼𝑟𝑎𝑑, the 𝛼𝑒𝑑𝑑𝑦 contribution should also be calculated and extracted from 𝛼𝑡𝑜𝑡𝑎𝑙 to extract the reduced damping parameter. 𝛼𝑒𝑑𝑑𝑦 can be estimated by the expression23 𝛼𝑒𝑑𝑑𝑦 =𝐶𝛾µ02𝑀𝑠𝛿2 16𝜌 , (6) where 𝐶 is a parameter describing the distribution of eddy current s within the films and its value is 0.5 in our studied samples and 𝜌 is the resistivity of the films. Resistivity is measured for all films with different dopant concentrations at different temperatures. It is in the range of 8.2×10-8 to 5.6 ×10-8 𝛺𝑚 for undoped, 5.7 ×10-7 to 5.3 ×10-7 𝛺𝑚 for 3.0 at% doped , 6.9 ×10-7 to 6.1 ×10- 7 𝛺𝑚 for 6.6 at% doped and 3.9×10-7 to 3.6 ×10-7 𝛺𝑚 for 12.6 at% doped films. Temperature dependent eddy current damping contributions , which are negligible, for all Fe 65Co35 films are given in Table 5. 𝛼𝑡𝑜𝑡 (filled symbols) and 𝛼𝑟𝑒𝑑 (open symbols) versus temperature for the differently Re -doped Fe65Co35 films are shown in Fig. 6 . Both damping parameter s slowly decrease with decreasing tempera ture. Moreover, the damping parameter increases with increasing Re concentration; the damping parameter is 4 times as large for the 12.6 at% Re -doped sample compared to the undoped sample . Since the damping parameter depends both on disorder induced scattering and spin-orbit coupling, the observed enhanc ement of the damping parameter can emerge from the electronic structure of the alloy and large spin -orbit coupling of Re. A c omparison between temperature dependent experimental 𝛼𝑡𝑜𝑡 and 𝛼𝑟𝑒𝑑 values and theoretically calculated intrinsic damping parameters is shown in Fig. 7 for the undoped and 12.6 at% Re -doped Fe 65Co35 films. In agreement with the experimental results, the theoretically calculated damping parameters decrease in magnitude with decreasing temperature . It has been argued by Schoen et al., 42 that the contribution to the intrinsic Gilbert damping parameter comes primarily from the strong electron -phonon coupling at high temperatures due to interband transition whereas at a low temperature, density of states at Fermi level (𝑛(𝐸𝐹)) and spin -orbit coupling give rise to intraband transition. In Fig. 8, we show the correspondence between the calculated damping parameter at 10 K with the density of states (spin up +spin down) at Fermi level for varying Re concentration. The increasing trend in both properties is obviously seen. The increase in DOS mainly comes from increasing DOS at Re sites in the spin -up channel. In the inset, the calculated spin -polarization as a function of Re concentration is shown. Spin polarization is defined as 𝜁=𝑛(𝐸𝐹)↑−𝑛(𝐸𝐹)↓ 𝑛(𝐸𝐹)↑+𝑛(𝐸𝐹)↓ where the contribution from both spin channels are seen. It is clearly observed that Re doping decreases the spin polarization. One should note that a quantitative comparison between theory and experiment requires more rigoro us theoretical considerations. The difference between experimental and theoretical results for the damping parameter may be explained by the incompleteness of the model used to calculate the Gilbert damping parameter by neglecting several complex scatterin g processes. Firstly, the effect of spin fluctuations was neglected, which in principle could be considered in the present methodology if the temperature dependent magnetization and hence information about the fluctuations of atomic moments were available from Monte -Carlo simulations. Other effects such as non-local damping and more sophisticated treatment of atomic displac ements in terms of phonon self -energies40 that may contribute to the relaxation of the magnetization in magnetic thin film materials have been neglected . Nevertheless, a qualitative agreement has been achieved where both experimental and theoretical results show that there is a significant increase of the damping parameter with increasing concentration of Re. 4. CONCLUSION Static and dynamic magnetic properties of rhenium doped Fe 65Co35 thin films have been investigated and clarified in a combined experimental and theoretical study. Results from first principles theoretical calculations show that the saturation magnetization gradually decreases with increasing Re concentration, from 2.3T for the undoped sample to 1.95T for the 10% Re -doped sample. The experimental results for the dependence of the saturation magnetization on the Re - doping are in agreement with the theoretical results, although indicating a more pronounced decrease of the saturation magnetization for the largest doping concentrations. The theoretical calculations show that the intrinsic Gilbert damping increases with increasing Re concentration; at room temperature the damping parameter is 2.8×10−3, which increases to 7.3×10−3 for the 10 at% Re -doped sample. Moreover, temperature dependent calculations of the Gilbert damping parameter reveal a weak decrease of the value with decreasing temperature . At a low temperature, our theoretical analysis showed the prominence of intra band contribution arising from an increase in the density of states at Fermi level. The experimental results for the damping parameter were corrected for radiative and eddy current contributions to the measured damping parameter and reveal similar trends as observed in the theoretical results; the damping parameter increases with increasing Re concentration and the damping parameter value decreases with decreasing temperature. The room temperature value for the reduced damping paramet er was 2.7×10−3 for the undoped sample, which increased to 9.0×10−3 for the 1 2.6 at% Re -doped film. The possibility to e nhanc e the damping parameter for Fe65Co35 thin films is a promising result since these materials are used in magnetic memory applications and higher data rates are achievable if the damping parameter of the material is increased. ACKNOWLEDGEMENT This work is supported by the Knut and Alice Wallenberg (KAW) Fou ndation, Grant No. KAW 2012.0031 and by the Marie Curie Action “Industry -Academia Partnership and Pathways” (ref. 612170, FP7 -PEOPLE -2013 -IAPP). The authors acknowledge financial support from Swedish Research Council (grant no. 2016 -05366) and Carl Tryggers Stiftelse (grant no. CTS 12:419 and 13:413). The simulations were performed on resources provided by the Swedish National Infrastructure (SNIC) at National Supercomputer Centre at Link öping University (NSC). M. Burak Kaynar is also acknowledged f or performing resistivity measurements. REFERENCES 1 M. Takagishi, H.N. Fuke, S. Hashimoto, H. Iwasaki, S. Kawasaki, R. Shiozaki, and M. Sahashi, J. Appl. Phys. 105, 07B725 (2009). 2 H. Fukuzawa, K. Koi, H. Tomita, H.N. Fuke, H. Iwasaki, and M. Sahashi, J. Appl. Phys. 91, 6684 (2002). 3 N. Nishimura, T. Hirai, A. Koganei, T. Ikeda, K. Okano, Y. Sekiguchi, and Y. Osada, J. Appl. Phys. 91, 5246 (2002). 4 H. Ohmori, T. Hatori, and S. Nakagawa, J. Magn. Magn. Mater. 320, 2963 (2008). 5 T. Nozaki, Y. Shiota, M. Shiraishi, T. Shinjo, and Y. Suzuki, Appl. Phys. Lett. 96, 022506 (2010). 6 N. Miyamoto, H. Ohmori, K. Mamiya, and S. Nakagawa, J. Appl. Phys. 107, 09C719 (2010). 7 N.X. Sun and S.X. Wang, IEEE Trans. Magn. 36, 2506 (2000). 8 S. Mizukami, E.P. Sajitha, D. Watanabe, F. Wu, T. Miyazaki, H. Naganuma, M. Oogane, and Y. Ando, Appl. Phys. Lett. 96, 152502 (2010). 9 H.S. Song, K.D. Lee, J.W. Sohn, S.H. Yang, S.S.P. Parkin, C.Y. You, and S.C. Shin, Appl. Phys. Lett. 102, 102401 (2013). 10 D. Lacour, J.A. Katine, N. Smith, M.J. Carey, and J.R. Childress, Appl. Phys. Lett. 85, 4681 (2004). 11 M. Ogiwara, S. Iihama, T. Seki, T. Kojima, S. Mizukami, M. Mizuguchi, a nd K. Takanashi, Appl. Phys. Lett. 103, 242409 (2013). 12 S. X. Wang, N. X. Sun, M. Yamaguchi, S. Yabukami, Nature 407, 150 (2000). 13 S. Ohnuma, H. Fujimori, T. Masumoto, X.Y. Xiong, D.H. Ping, and K. Hono, Appl. Phys. Lett. 82, 946 (2003). 14 S. Ohnuma, N. Kobayashi, H. Fujimori, T. Masumoto, X.Y. Xiong, and K. Hono, Scr. Mater. 48, 903 (2003). 15 A. Hashimoto, K. Hirata, T. Matsuu, S. Saito, and S. Nakagawa, IEEE Trans. Magn. 44, 3899 (2008). 16 Y. Fu, T. Miyao, J.W. Cao, Z. Yang, M. Matsumot o, X.X. Liu, and A. Morisako, J. Magn. Magn. Mater. 308, 165 (2007). 17 M.C. Hickey and J.S. Moodera, Phys. Rev. Lett. 102, 137601 (2009). 18 S. Ingvarsson, L. Ritchie, X.Y. Liu, G. Xiao, J.C. Slonczewski, P.L. Trouilloud, and R.H. Koch, Phys. Rev. B 66, 214416 (2002). 19 G. Counil, T. Devolder, J. V. Kim, P. Crozat, C. Chappert, S. Zoll, and R. Fournel, IEEE Trans. Magn. 42, 3323 (2006). 20 K. Wu, M. Tang, D. Li, X. Guo, B. Cui, J. Yun, Y. Zuo, L. Xi, and Z.Z. Zhang, J. Phys. D. Appl. Phys. 50, 135001 (20 17). 21 I. Barsukov, P. Landeros, R. Meckenstock, J. Lindner, D. Spoddig, Z.A. Li, B. Krumme, H. Wende, D.L. Mills, and M. Farle, Phys. Rev. B 85, 014420 (2012). 22 J. Lindner, I. Barsukov, C. Raeder, C. Hassel, O. Posth, R. Meckenstock, P. Landeros, and D.L. Mills, Phys. Rev. B 80, 224421 (2009). 23 M.A.W. Schoen, J.M. Shaw, H.T. Nembach, M. Weiler, and T.J. Silva, Phys. Rev. B 92, 184417 (2015). 24 Y. Li and W.E. Bailey, Phys. Rev. Lett. 116, 117602 (2016). 25 I. Radu, G. Woltersdorf, M. Kiessling, A. Melnikov, U. Bovensiepen, J.U. Thiele, and C.H. Back, Phys. Rev. Lett. 102, 117201 (2009). 26 S.E. Russek, P. Kabos, R.D. McMichael, C.G. Lee, W.E. Bailey, R. Ewasko, and S.C. Sanders, J. Appl. Phys. 91, 8659 (2002). 27 W. Bailey, P. Kabos, F. Mancoff, and S. Russek, IEEE Trans. Magn. 37, 1749 (2001). 28 S. Ingvarsson, G. Xiao, S.S.P. Parkin, and R.H. Koch, Appl. Phys. Lett. 85, 4995 (2004). 29 J.O. Rantschler, R.D. McMichael, A. Castillo, A.J. Shapiro, W.F. Eg elhoff, B.B. Maranville, D. Pulugurtha, A.P. Chen, and L.M. Connors, J. Appl. Phys. 101, 033911 (2007). 30 F. Pan, Gilbert Damping of Doped Permalloy from First Principles Calculations , Licenciate Thesis in Physics, KTH Royal Institute of Technology , Swe den (2015). 31 F. Xu, Z. Xu, and Y. Yin, IEEE Trans. Magn. 51, 2800904 (2015). 32 L. Sun, G.N. Sun, J.S. Wang, P.K.J. Wong, Y. Zhang, Y. Fu, Y. Zhai, X.J. Bai, J. Du, and H.R. Zhai, J. Nanosci. Nanotechnol. 12, 6562 (2012). 33 F. Xu, X. Chen, Y. Ma, N.N. Phuoc, X. Zhang, and C.K. Ong, J. Appl. Phys. 104, 083915 (2008). 34 W. Zhang, D. Zhang, P.K.J. Wong, H. Yuan, S. Jiang, G. Van Der Laan, Y. Zhai, and Z. Lu, ACS Appl. Mater. Interfaces 7, 17070 (2015). 35 R. Banerjee, C. Autie ri, V. A. Venugopal, S. Basu, M. A. Gubbins and B. Sanyal, unpublished. 36 W. K. Chu, J. W. Mayer, and M. -A. Nicolet, Backscattering Spectrometry, Academic Press , Orlando , 1978 . 37 M. Mayer, SIMNRA User's Guide , Report IPP 9/113, Max -Planck -Institut für Plasmaphysik, Garching, Germany (1997) 38 S. Mankovsky, D. Ködderitzsch, G. Woltersdorf, and H. Ebert, Phys. Rev. B 87, 014430 (2013). 39 W. Wang, Y. Chen, G.H. Yue, K. Sumiyama, T. Hihara, and D.L. Peng, J. Appl. Phys. 106, 013912 (2009). 40 A. Kumar, F. Pan, S. Husain, S. Akansel, R. Brucas, L. Bergqvist, S. Chaudhary, and P. Svedlindh, Phys. Rev. B 96, 224425 (2017). 41 H.T. Nembach, T.J. Silva, J.M. Shaw, M.L. Schneider, M.J. Carey, S. Maat, and J.R. Childress, Phys. Rev. B 84, 054424 (2011). 42 M. A. W. Schoen, D. Thonig, M. L. Schneider, T. J. Silva, H. T. Nembach, O. Eriksson, O. Karis, and J. M. Shaw, Nature Physics, 12, 839 –842 (2016). Figure 1 (a) GIXRD plot for Fe 65Co35 films with dif ferent Re concentrations. S hift of (110) peak diffraction peak with Re concentration is given as insert . (b) Lattice parameter versus Re concentration. Circles are lattice parameters extracted from XRD measurements and squares are calculated th eoretical values. Line s are guide to the eye. Figure 2 (a) Normalized room temperature magnetization versus magnetic field for Fe 65Co35 films with different Re concentration . (b) Low temperature saturation magnetization versus Re concentration. Circles are experimental data and squares corresponding calculated results. Experimental 𝝁𝟎𝑴𝒔 values were extracted from temperature dependent FMR results. Lines are guide s to the eye . Figure 3 𝝁𝟎𝑯𝒓 versus in -plane angle of magnetic field 𝝓𝑯 for different dopant concentrations of Re. Black squares are experimental data and red line s are fits to Eq. (1). Figure 4 Room temperature real (a and c) and imaginary (b and d) 𝑺𝟐𝟏 components versus out - of-plane magnetic field for Fe65Co35 thin films with 0% and 12.6 at% Re recorded at 20GHz . Black squares are data points and red lines are fit s to Eq. (2). Figure 5 (a) Frequency versus 𝝁𝟎𝑯𝒓 values at different temperatures for the Fe65Co35 thin film with 12.6 at% Re. Coloured lines correspond to fits to Eq. ( 4). (b) Linewidth 𝝁𝟎∆𝑯 versus frequency at different temperatures for the same Re doping concentration. Coloured lines correspond to fits to Eq. ( 3). Symbols represent experimental data. Figur e 6 𝜶𝒕𝒐𝒕 versus temperature for Fe 65Co35 thin films with different concentration of Re. Besides showing 𝜶𝒕𝒐𝒕 , reduced 𝜶𝒓𝒆𝒅 values are also plotted obtained by subtraction of radiative damping and eddy current damping contributions from 𝜶𝒕𝒐𝒕. Error bars are given for measured 𝜶𝒕𝒐𝒕 (same size as symbol size ). Figur e 7 𝜶𝒕𝒐𝒕 versus temperature for Fe 65Co35 thin films with 0 at% and 1 2.6 at% concentration of Re. Beside s showing 𝜶𝒕𝒐𝒕 , reduced 𝜶𝒓𝒆𝒅 values are also plotted obtained by subtraction of radiative damping and eddy current damping contribution s from 𝜶𝒕𝒐𝒕. In addition to experimental results theoretically calculated intrinsic damping parameters are given for the similar concentrations of Re . Error bars are given for measured 𝜶𝒕𝒐𝒕 (same size as symbol size) . Figure 8 Calculated density of states at Fermi level (left axis) and damping parameter (right axis) are shown as a function of Re concentration. In the inset, spin -polarization is shown as a function of Re concentration. 0 0.03 0.06 0.09 0.12 Re concentration0.90.951DOS at EF (States/eV) 0 0.03 0.06 0.09 0.12 Re concentration0.350.40.450.50.55Spin polarization 0123456 Damping parameter (x 10-3)Re (at%) 𝑡𝑅𝑢,𝑐𝑎𝑝 (nm) 𝜎 (nm) 𝑡𝐹𝑒𝐶𝑜 (nm) 𝜎 (nm) 𝑡𝑅𝑢,𝑠𝑒𝑒𝑑 (nm) (nm) 0 2.46 1.89 39.71 0.67 2.74 0.66 3.0 2.47 1.80 37.47 0.59 2.45 1.03 6.6 1.85 0.50 37.47 0.51 2.13 0.90 12.6 2.15 1.49 37.38 0.64 1.89 1.03 Table 1 Thickness and roughness (𝝈) values for different layers in films extracted from XRR data. Error margin is 0.02nm for all thickness and roughness values. Re (at%) 𝜇0𝐻𝑢 (mT) 𝜇0𝑀𝑒𝑓𝑓 (T) 0 2.20 2.31 3.0 2.10 2.12 6.6 2.30 1.95 12.6 2.20 1.64 Table 2 Room temperature 𝝁𝟎𝑴𝒆𝒇𝒇 and 𝝁𝟎𝑯𝒖 values for Fe 65Co35 films with different concentration of Re extracted by fitting the angle dependent cavity FMR data to Eq. (1). Temperature (K) 0% Re 3.0 at% Re 6.6 at% Re 12.6 at% Re 𝜇0𝑀𝑒𝑓𝑓 (T) 𝜇0𝑀𝑒𝑓𝑓 (T) 𝜇0𝑀𝑒𝑓𝑓 (T) 𝜇0𝑀𝑒𝑓𝑓 (T) 300 2.29 2.16 1.99 1.61 200 2.31 2.16 2.04 1.67 150 2.33 2.24 2.06 1.70 100 2.36 2.25 2.07 1.72 50 2.36 2.27 2.08 1.74 Table 3 Temperature dependent 𝝁𝟎𝑴𝒆𝒇𝒇 values for Fe65Co35 films with different concentrati on of Re extracted by fitting broadband out -of-plane FMR data to Eq. (4). Error margin is about 10 mT. Temperature(K) 𝛼𝑟𝑎𝑑 (×10-3) 0% Re 3.0 at% Re 6.6 at% Re 12.6 at% Re 300 0.218 0.482 0.438 0.154 200 0.222 0.494 0.450 0.160 150 0.216 0.499 0.454 0.162 100 0.225 0.502 0.456 0.219 50 0.221 0.505 0.458 0.166 Table 4 Temperature dependent r adiative damping contribution to total damping parameter for Fe65Co35 films with different concentration of Re calculated using Eq. (5). Temperature(K) 𝛼𝑒𝑑𝑑𝑦 (×10-3) 0% Re 3.3 at% Re 6.6 at% Re 12.6 at% Re 300 0.038 0.077 0.064 0.006 200 0.047 0.081 0.067 0.006 150 0.050 0.084 0.070 0.006 100 0.055 0.084 0.073 0.007 50 0.058 0.086 0.075 0.007 Table 5 Temperature dependent eddy current damping contribution to total damping parameter for Fe 65Co35 films with different concentration of Re calculated using Eq. ( 6). Figure A1 RBS spectra for the Re -doped Fe 65Co35 films. | 2019-02-26 | The effects of rhenium doping in the range 0 to 10 atomic percent on the
static and dynamic magnetic properties of Fe65Co35 thin films have been studied
experimentally as well as with first principles electronic structure
calculations focusing on the change of the saturation magnetization and the
Gilbert damping parameter. Both experimental and theoretical results show that
the saturation magnetization decreases with increasing Re doping level, while
at the same time Gilbert damping parameter increases. The experimental low
temperature saturation magnetic induction exhibits a 29 percent decrease, from
2.31 T to 1.64 T, in the investigated doping concentration range, which is more
than predicted by the theoretical calculations. The room temperature value of
the damping parameter obtained from ferromagnetic resonance measurements,
correcting for extrinsic contributions to the damping, is for the undoped
sample 0.0027, which is close to the theoretically calculated Gilbert damping
parameter. With 10 atomic percent Re doping, the damping parameter increases to
0.0090, which is in good agreement with the theoretical value of 0.0073. The
increase in damping parameter with Re doping is explained by the increase in
density of states at Fermi level, mostly contributed by the spin-up channel of
Re. Moreover, both experimental and theoretical values for the damping
parameter are observed to be weakly decreasing with decreasing temperature. | Enhanced Gilbert Damping in Re doped FeCo Films: A Combined Experimental and Theoretical Study | 1902.09896v1 |
arXiv:1408.3499v1 [math.AP] 15 Aug 2014Linear hyperbolic equations with time-dependent propagation speed and strong damping Marina Ghisi Universit` a degli Studi di Pisa Dipartimento di Matematica PISA (Italy) e-mail:ghisi@dm.unipi.itMassimo Gobbino Universit` a degli Studi di Pisa Dipartimento di Matematica PISA (Italy) e-mail:m.gobbino@dma.unipi.itAbstract We consider a second order linear equation with a time-dependent co efficientc(t) in front of the “elastic” operator. For these equations it is well-know n that a higher space- regularity of initial data compensates a lower time-regularity of c(t). In this paper we investigate the influence of a strong dissipation, na mely a friction term which depends on a power of the elastic operator. What we discover is a threshold effect. When the exponent of the ela stic operator in the friction term is greater than 1/2, the damping prevails and the equation behaves as if the coefficient c(t) were constant. When the exponent is less than 1/2, the time- regularity of c(t) comes into play. If c(t) is regular enough, once again the damping prevails. On the contrary, when c(t) is not regular enough the damping might be ineffective, and there are examples in which the dissipative equation b ehaves as the non-dissipative one. As expected, the stronger is the damping, th e lower is the time- regularity threshold. We also provide counterexamples showing the optimality of our result s. Mathematics Subject Classification 2010 (MSC2010): 35L20, 35L80, 35L90. Key words: linear hyperbolic equation, dissipative hyperbolic equation, strong d amp- ing, fractional damping, time-dependent coefficients, well-posedn ess, Gevrey spaces.1 Introduction LetHbe a separable real Hilbert space. For every xandyinH,|x|denotes the norm ofx, and/a\}⌊ra⌋ketle{tx,y/a\}⌊ra⌋ketri}htdenotes the scalar product of xandy. LetAbe a self-adjoint linear operator on Hwith dense domain D(A). We assume that Ais nonnegative, namely /a\}⌊ra⌋ketle{tAx,x/a\}⌊ra⌋ketri}ht ≥0 for every x∈D(A), so that for every α≥0 the power Aαxis defined provided that xlies in a suitable domain D(Aα). We consider the second order linear evolution equation u′′(t)+2δAσu′(t)+c(t)Au(t) = 0, (1.1) with initial data u(0) =u0, u′(0) =u1. (1.2) As far as we know, this equation has been considered in the literatur e either in the case where δ= 0, or in the case where δ >0 but the coefficient c(t) is constant. Let us give a brief outline of the previous literature which is closely related to our results. The non-dissipative case Whenδ= 0, equation (1.1) reduces to u′′(t)+c(t)Au(t) = 0. (1.3) This is the abstract setting of a wave equation in which c(t) represents the square of the propagation speed. If the coefficient c(t) is Lipschitz continuous and satisfies the strict hyperbolicity condition 0<µ1≤c(t)≤µ2, (1.4) then it is well-know that problem (1.3)–(1.2) is well-posed in the classic e nergy space D(A1/2)×H(see for example the classic reference [14]). If the coefficient is not Lipschitz continuous, things are more comple x, even if (1.4) still holds true. This problem was addressed by F. Colombini, E. De Gior gi and S. Spag- nolo in the seminal paper [6]. Their results can be summed up as follows ( we refer to section 2 below for the precise functional setting and rigorous sta tements). (1) Problem (1.3)–(1.2) has always a unique solution, up to admitting t hat this solu- tion takes its values in a very large Hilbert space (ultradistributions) . This is true for initial data in the energy space D(A1/2)×H, but also for less regular data, such as distributions or ultradistributions. (2) If initial data are regular enough, then the solution is regular as well. How much regularity is required depends on the time-regularity of c(t). Classic examples are the following. If c(t) is just measurable, problem (1.3)–(1.2) is well-posed in the class of analytic functions. If c(t) isα-H¨ older continuous for some α∈(0,1), problem (1.3)–(1.2) is well-posed in the Gevrey space of order (1 −α)−1. 1(3) If initial data are not regular enough, then the solution may exh ibit a severe derivative loss for all positive times. For example, for every α∈(0,1) there exist a coefficientc(t) which isα-H¨ older continuous, and initial data ( u0,u1) which are in the Gevrey class of order βfor everyβ >(1−α)−1, such that the corresponding solution to (1.3)–(1.2) (which exists in the weak sense of point (1)) is not even a distribution for every t>0. In the sequel we call (DGCS)-phenomenon the instantaneous loss of regularity de- scribed in point (3) above. The dissipative case with constant coefficients Ifδ >0 andc(t) is a constant function (equal to 1 without loss of generality), equation (1.1) reduces to u′′(t)+2δAσu′(t)+Au(t) = 0. (1.5) Mathematical models with damping terms of this form were proposed in [1], and then rigorously analyzed by many authors from different points of v iew. The first papers [2, 3, 4], and the more recent [10], are devoted to analyticity properties of the semigroup associated to (1.5). The classic assumptions in these pap ers are that the operatorAis strictly positive, σ∈[0,1], and the phase space is D(A1/2)×H. On a different side, the community working on dispersive equations consid ered equation (1.5) intheconcretecasewhere σ∈[0,1]andAu=−∆uinRnorspecialclassesofunbounded domains. They proved energy decay and dispersive estimates, but exploiting in an essential way the spectral properties of the Laplacian in those do mains. The interested reader is referred to [11, 12, 13, 19] and to the references quot ed therein. Finally, equation (1.5) was considered in [9] in full generality, namely fo r every σ≥0 and every nonnegative self-adjoint operator A. Two different regimes appeared. In the subcritical regime σ∈[0,1/2], problem (1.5)–(1.2) is well-posed in the classic energy space D(A1/2)×Hor more generally in D(Aα+1/2)×D(Aα) withα≥0. In the supercritical regime σ≥1/2, problem (1.5)–(1.2) is well-posed in D(Aα)×D(Aβ) if and only if 1−σ≤α−β≤σ. (1.6) This means that in the supercritical regime different choices of the p hase space are possible, even with α−β/\e}atio\slash= 1/2. The dissipative case with time-dependent coefficients As far as we know, the case of a dissipative equation with a time-dependent propagation speed had n ot been considered yet. The main question we address in this paper is the extent to which the dissipative term added in (1.1) prevents the (DGCS)-phenomenon of (1.3) fro m happening. We discover a composite picture, depending on σ. •In the subcritical regime σ∈[0,1/2], if the strict hyperbolicity assumption (1.4) is satisfied, well-posedness results do depend on the time-regularit y ofc(t) (see Theorem 3.2). Classic examples are the following. 2–Ifc(t) isα-H¨ older continuous for some exponent α >1−2σ, then the dis- sipation prevails, and problem (1.1)–(1.2) is well-posed in the classic en ergy spaceD(A1/2)×Hor more generally in D(Aβ+1/2)×D(Aβ) withβ≥0. –Ifc(t) is no more than α-H¨ older continuous for some exponent α <1−2σ, thenthedissipationcanbeneglected, sothat(1.1)behavesexact lyasthenon- dissipative equation (1.3). This means well-posedness in the Gevrey s pace of order (1−α)−1and the possibility to produce the (DGCS)-phenomenon for less regular data (see Theorem 3.10). –The case with α= 1−2σis critical and also the size of the H¨ older constant ofc(t) compared with δcomes into play. •In the supercritical regime σ >1/2 the dissipation prevails in an overwhelming way. In Theorem 3.1 we prove that, if c(t) is just measurable and satisfies just the degenerate hyperbolicity condition 0≤c(t)≤µ2, (1.7) then (1.1) behaves as (1.5). This means that problem (1.1)–(1.2) is w ell-posed in D(Aα)×D(Aβ) if and only if (1.6) is satisfied, the same result obtained in [9] in the case of a constant coefficient. The second issue we address in this paper is the further space-reg ularity of solutions for positive times, since a strong dissipation is expected to have a re gularizing effect similar to parabolic equations. This turns out to be true provided tha t the assumptions of our well-posedness results are satisfied, and in addition σ∈(0,1). Indeed, we prove that in this regime u(t) lies in the Gevrey space of order (2min {σ,1−σ})−1for every t>0. We refer to Theorem 3.8 and Theorem 3.9 for the details. This effec t had already been observed in [15] in the dispersive case. We point out that the regularizing effect is maximum when σ= 1/2 (the only case in which solutions become analytic with respect to space variables) and disappears when σ≥1, meaning that a stronger overdamping prevents smoothing. Overview of the technique The spectral theory reduces the problem to an analysis of the family of ordinary differential equations u′′ λ(t)+2δλ2σu′ λ(t)+λ2c(t)uλ(t) = 0. (1.8) Whenδ= 0, a coefficient c(t) which oscillates with a suitable period can produce a resonance effect so that (1.8) admits a solution whose oscillations h ave an amplitude which grows exponentially with time. This is the primordial origin of the ( DGCS)- phenomenon for non-dissipative equations. When δ >0, the damping term causes an exponential decay of the amplitude of oscillations. The competition between the 3exponential energy growth due to resonance and the exponent ial energy decay due to dissipation originates the threshold effect we observed. Whenc(t) is constant, equation (1.8) can be explicitly integrated, and the ex plicit formulae for solutions led to the sharp results of [9]. Here we need th e same sharp estimates, but without relying on explicit solutions. To this end, we int roduce suitable energy estimates. In the supercritical regime σ≥1/2 we exploit the following σ-adapted “Kovaleskyan energy” E(t) :=|u′ λ(t)+δλ2σuλ(t)|2+δ2λ4σ|uλ(t)|2. (1.9) In the subcritical regime σ≤1/2 we exploit the so-called “approximated hyperbolic energies” Eε(t) :=|u′ λ(t)+δλ2σuλ(t)|2+δ2λ4σ|uλ(t)|2+λ2cε(t)|uλ(t)|2,(1.10) obtained by adding to (1.9) an “hyperbolic term” depending on a suita ble smooth ap- proximation cε(t) ofc(t), which in turn is chosen in a λ-dependent way. Terms of this type are the key tool introduced in [6] for the non-dissipative equa tion. Future extensions We hope that this paper could represent a first step in the theory of dissipative hyperbolic equations with variable coefficients, both line ar and nonlinear. Next steps could be considering a coefficient c(x,t) depending both on time and space variables, and finally quasilinear equations. This could lead to improve t he classic results by K. Nishihara [16, 17] for Kirchhoff equations, whose linear ization has a time- dependent coefficient, and finally to consider more general local no nlinearities, in which case the linearization involves a coefficient c(x,t) depending on both variables. Inadifferent direction, thesubcritical case σ∈[0,1/2]withdegeneratehyperbolicity assumptions remains open and could be the subject of further res earch, in the same way as [7] was the follow-up of [6]. On the other side, we hope that our counterexamples could finally dis pel the dif- fuse misconception according to which dissipative hyperbolic equatio ns are more stable, and hence definitely easier to handle. Now we know that a friction ter m below a suit- able threshold is substantially ineffective, opening the door to patho logies such as the (DGCS)-phenomenon, exactly as in the non-dissipative case. Structure of the paper This paper is organized as follows. In section 2 we introduce the functional setting and we recall the classic existence results f rom [6]. In section 3 we state our main results. In section 4 we provide a heuristic descriptio n of the competition between resonance and decay. In section 5 we prove our existenc e and regularity results. In section 6 we present our examples of (DGCS)-phenomenon. 42 Notation and previous results Functional spaces LetHbe a separable Hilbert space. Let us assume that Hadmits a countable complete orthonormal system {ek}k∈Nmade by eigenvectors of A. We denote the corresponding eigenvalues by λ2 k(with the agreement that λk≥0), so that Aek=λ2 kekfor everyk∈N. In this case every u∈Hcan be written in a unique way in the form u=/summationtext∞ k=0ukek, whereuk=/a\}⌊ra⌋ketle{tu,ek/a\}⌊ra⌋ketri}htare the Fourier components of u. In other words, the Hilbert space Hcan be identified with the set of sequences {uk}of real numbers such that/summationtext∞ k=0u2 k<+∞. We stress that this is just a simplifying assumption, with substantially no loss of generality. Indeed, according to the spectral theorem in its gene ral form (see for ex- ample Theorem VIII.4 in [18]), one can always identify HwithL2(M,µ) for a suitable measure space ( M,µ), in such a way that under this identification the operator Aacts as a multiplication operator by some measurable function λ2(ξ). All definitions and statements in the sequel, with the exception of the counterexamp les of Theorem 3.10, can be easily extended to the general setting just by replacing the sequence {λ2 k}with the function λ2(ξ), and the sequence {uk}of Fourier components of uwith the element /hatwideu(ξ) ofL2(M,µ) corresponding to uunder the identification of HwithL2(M,µ). The usual functional spaces can be characterized in terms of Fou rier components as follows. Definition 2.1. Letube a sequence {uk}of real numbers. •Sobolev spaces . For every α≥0 it turns out that u∈D(Aα) if /⌊ard⌊lu/⌊ard⌊l2 D(Aα):=∞/summationdisplay k=0(1+λk)4αu2 k<+∞. (2.1) •Distributions . We say that u∈D(A−α) for someα≥0 if /⌊ard⌊lu/⌊ard⌊l2 D(A−α):=∞/summationdisplay k=0(1+λk)−4αu2 k<+∞. (2.2) •Generalized Gevrey spaces . Letϕ: [0,+∞)→[0,+∞) be any function, let r≥0, and letα∈R. We say that u∈ Gϕ,r,α(A) if /⌊ard⌊lu/⌊ard⌊l2 ϕ,r,α:=∞/summationdisplay k=0(1+λk)4αu2 kexp/parenleftbig 2rϕ(λk)/parenrightbig <+∞. (2.3) •Generalized Gevrey ultradistributions . Letψ: [0,+∞)→[0,+∞)beanyfunction, letR≥0, and letα∈R. We say that u∈ G−ψ,R,α(A) if /⌊ard⌊lu/⌊ard⌊l2 −ψ,R,α:=∞/summationdisplay k=0(1+λk)4αu2 kexp/parenleftbig −2Rψ(λk)/parenrightbig <+∞. (2.4) 5Remark 2.2. Ifϕ1(x) =ϕ2(x) for every x >0, thenGϕ1,r,α(A) =Gϕ2,r,α(A) for every admissible value of randα. For this reason, with a little abuse of notation, we consider the spaces Gϕ,r,α(A) even when ϕ(x) is defined only for x >0. The same comment applies also to the spaces G−ψ,R,α(A). The quantities defined in (2.1) through (2.4) are actually norms which induce a Hilbert space structure on D(Aα),Gϕ,r,α(A),G−ψ,R,α(A), respectively. The standard inclusions Gϕ,r,α(A)⊆D(Aα)⊆H⊆D(A−α)⊆ G−ψ,R,−α(A) hold true for every α≥0 and every admissible choice of ϕ,ψ,r,R. All inclusions are strict if α,randRare positive, and the sequences {λk},{ϕ(λk)}, and{ψ(λk)}are unbounded. We observe that Gϕ,r,α(A) is actually a so-called scale of Hilbert spaces with respect to theparameter r, withlarger values of rcorresponding to smaller spaces. Analogously, G−ψ,R,α(A) is a scale of Hilbert spaces with respect to the parameter R, but with larger values ofRcorresponding to larger spaces. Remark 2.3. Let us consider the concrete case where I⊆Ris an open interval, H=L2(I), andAu=−uxx, with periodic boundary conditions. For every α≥0, the spaceD(Aα) is actually the usual Sobolev space H2α(I), andD(A−α) is the usual space of distributions of order 2 α. Whenϕ(x) :=x1/sfor somes>0, elements of Gϕ,r,0(A) withr>0 are usually called Gevrey functions of order s, the cases= 1 corresponding to analytic functions. When ψ(x) :=x1/sforsomes>0, elements of G−ψ,R,0(A)withR>0areusually called Gevrey ultradistributions of order s, the cases= 1 corresponding to analytic functionals. In this case the parameter αis substantially irrelevant because the exponential term is dominant both in (2.3) and in (2.4). For the sake of consistency, with a little abuse of notation we use th e same terms (Gevrey functions, Gevrey ultradistributions, analytic functions and analytic function- als) in order to denote the same spaces also in the general abstrac t framework. To be more precise, we should always add “with respect to the operator A”, or even better “with respect to the operator A1/2”. Continuity moduli Throughout this paper we call continuity modulus any continuous functionω: [0,+∞)→[0,+∞) such that ω(0) = 0,ω(x)>0 for every x >0, and moreover x→ω(x) is a nondecreasing function , (2.5) x→x ω(x)is a nondecreasing function. (2.6) A function c: [0,+∞)→Ris said to be ω-continuous if |c(a)−c(b)| ≤ω(|a−b|)∀a≥0,∀b≥0. (2.7) 6More generally, a function c:X→R(withX⊆R) is said to be ω-continuous if it satisfies the same inequality for every aandbinX. Previous results We are now ready to recall the classic results concerning existence , uniqueness, and regularity for solutions to problem (1.1)–(1.2). We state them using our notations which allow general continuity moduli and general spaces of Gevrey functions or ultradistributions. Proofs are a straightforward application of the approximated ene rgy estimates in- troduced in [6]. In that paper only the case δ= 0 is considered, but when δ≥0 all new terms have the “right sign” in those estimates. The first result concerns existence and uniqueness in huge spaces such as analytic functionals, with minimal assumptions on c(t). Theorem A (see [6, Theorem 1]) .Let us consider problem (1.1)–(1.2) under the fol- lowing assumptions: •Ais a self-adjoint nonnegative operator on a separable Hilbe rt spaceH, •c∈L1((0,T))for everyT >0(without sign conditions), •σ≥0andδ≥0are two real numbers, •initial conditions satisfy (u0,u1)∈ G−ψ,R0,1/2(A)×G−ψ,R0,0(A) for someR0>0and someψ: (0,+∞)→(0,+∞)such that limsup x→+∞x ψ(x)<+∞. Then there exists a nondecreasing function R: [0,+∞)→[0,+∞), withR(0) =R0, such that problem (1.1)–(1.2) admits a unique solution u∈C0/parenleftbig [0,+∞);G−ψ,R(t),1/2(A)/parenrightbig ∩C1/parenleftbig [0,+∞);G−ψ,R(t),0(A)/parenrightbig .(2.8) Condition (2.8), with the range space increasing with time, simply mean s that u∈C0/parenleftbig [0,τ];G−ψ,R(τ),1/2(A)/parenrightbig ∩C1/parenleftbig [0,τ];G−ψ,R(τ),0(A)/parenrightbig ∀τ≥0. This amounts to say that scales of Hilbert spaces, rather than fixe d Hilbert spaces, are the natural setting for this problem. Inthesecondresultweassumestricthyperbolicityand ω-continuityofthecoefficient, and we obtain well-posedness in a suitable class of Gevrey ultradistrib utions. 7Theorem B (see [6, Theorem 3]) .Let us consider problem (1.1)–(1.2) under the fol- lowing assumptions: •Ais a self-adjoint nonnegative operator on a separable Hilbe rt spaceH, •the coefficient c: [0,+∞)→Rsatisfies the strict hyperbolicity assumption (1.4) and theω-continuity assumption (2.7) for some continuity modulus ω(x), •σ≥0andδ≥0are two real numbers, •initial conditions satisfy (u0,u1)∈ G−ψ,R0,1/2(A)×G−ψ,R0,0(A) for someR0>0and some function ψ: (0,+∞)→(0,+∞)such that limsup x→+∞x ψ(x)ω/parenleftbigg1 x/parenrightbigg <+∞. (2.9) Letube the unique solution to the problem provided by Theorem A. Then there exists R>0such that u∈C0/parenleftbig [0,+∞),G−ψ,R0+Rt,1/2(A)/parenrightbig ∩C1([0,+∞),G−ψ,R0+Rt,0(A)). The third result we recall concerns existence of regular solutions. The assumptions onc(t) are the same as in Theorem B, but initial data are significantly more r egular (Gevrey spaces instead of Gevrey ultradistributions). Theorem C (see [6, Theorem 2]) .Let us consider problem (1.1)–(1.2) under the fol- lowing assumptions: •Ais a self-adjoint nonnegative operator on a separable Hilbe rt spaceH, •the coefficient c: [0,+∞)→Rsatisfies the strict hyperbolicity assumption (1.4) and theω-continuity assumption (2.7) for some continuity modulus ω(x), •σ≥0andδ≥0are two real numbers, •initial conditions satisfy (u0,u1)∈ Gϕ,r0,1/2(A)×Gϕ,r0,0(A) for somer0>0and some function ϕ: (0,+∞)→(0,+∞)such that limsup x→+∞x ϕ(x)ω/parenleftbigg1 x/parenrightbigg <+∞. (2.10) 8Letube the unique solution to the problem provided by Theorem A. Then there exist T >0andr>0such thatrT <r 0and u∈C0/parenleftbig [0,T],Gϕ,r0−rt,1/2(A)/parenrightbig ∩C1([0,T],Gϕ,r0−rt,0(A)). (2.11) Remark 2.4. The key assumptions of Theorem B and Theorem C are (2.9) and (2.10 ), respectively, representing the exact compensation between spa ce-regularity of initial dataandtime-regularityofthecoefficient c(t)requiredinordertoobtainwell-posedness. These conditions do not appear explicitly in [6], where they are replace d by suitable specific choices of ω,ϕ,ψ, which of course satisfy the same relations. To our knowledge, those conditions were stated for the first time in [8], thus unifying se veral papers that in thelast 30 years hadbeendevoted tospecial cases (see forexam ple [5] andthe references quoted therein). Remark 2.5. The standard example of application of Theorem B and Theorem C is the following. Let us assume that c(t) isα-H¨ older continuous for some α∈(0,1), namelyω(x) =Mxαfor a suitable constant M. Then (2.9) and (2.10) hold true with ψ(x) =ϕ(x) :=x1−α. This leads to well-posedness both in the large space of Gevrey ultradistributions of order (1 −α)−1, and in the small space of Gevrey functions of the same order. Remark 2.6. The choice of ultradistributions in Theorem B is not motivated by the searchforgenerality, butitisinsomesense theonlypossibleonebec auseofthe(DGCS)- phenomenon exhibited in [6], at least in the non-dissipative case. When δ= 0, if initial data are taken in Sobolev spaces or in any space larger than the Gev rey spaces of Theorem C, then it may happen that for all positive times the solution lies in the space of ultradistributions specified in Theorem B, and nothing more. In ot her words, for δ= 0 there is no well-posedness result in between the Gevrey spaces o f Theorem C and the Gevrey ultradistributions of Theorem B, and conditions (2.9 ) and (2.10) are optimal. The aim of this paper is to provide an optimal picture for the case δ >0. 3 Main results In this section we state our main regularity results for solutions to ( 1.1)–(1.2). To this end, we need some further notation. Given any ν≥0, we write Has an orthogonal direct sum H:=Hν,−⊕Hν,+, (3.1) whereHν,−is the closure of the subspace generated by all eigenvectors of Arelative to eigenvalues λk<ν, andHν,+is the closure of the subspace generated by all eigenvectors ofArelative to eigenvalues λk≥ν. For every vector u∈H, we writeuν,−anduν,+ to denote its components with respect to the decomposition (3.1). We point out that 9Hν,−andHν,+areA-invariant subspaces of H, and thatAis a bounded operator when restricted to Hν,−, and a coercive operator when restricted to Hν,+ifν >0. In the following statements we provide separate estimates for low- frequency compo- nentsuν,−(t) and high-frequency components uν,+(t) of solutions to (1.1). This is due to the fact that the energy of uν,−(t) can be unbounded as t→+∞, while in many cases we are able to prove that the energy of uν,+(t) is bounded in time. 3.1 Existence results in Sobolev spaces The first result concerns the supercritical regime σ≥1/2, in which case the dissipation always dominates the time-dependent coefficient. Theorem 3.1 (Supercritical dissipation) .Let us consider problem (1.1)–(1.2) under the following assumptions: •Ais a self-adjoint nonnegative operator on a separable Hilbe rt spaceH, •the coefficient c: [0,+∞)→Ris measurable and satisfies the degenerate hyper- bolicity assumption (1.7), •σandδare two positive real numbers such that either σ >1/2, orσ= 1/2and 4δ2≥µ2, •(u0,u1)∈D(Aα)×D(Aβ)for some real numbers αandβsatisfying (1.6). Letube the unique solution to the problem provided by Theorem A. Thenuactually satisfies (u,u′)∈C0/parenleftbig [0,+∞),D(Aα)×D(Aβ)/parenrightbig . (3.2) Moreover, for every ν≥1such that 4δ2ν4σ−2≥µ2, it turns out that |Aβu′ ν,+(t)|2+|Aαuν,+(t)|2≤/parenleftbigg 2+2 δ2+µ2 2 δ4/parenrightbigg |Aβu1,ν,+|2+3/parenleftbigg 1+µ2 2 2δ2/parenrightbigg |Aαu0,ν,+|2(3.3) for everyt≥0. Our second result concerns the subcritical regime σ∈[0,1/2], in which case the time-regularity of c(t) competes with the exponent σ. Theorem 3.2 (Subcritical dissipation) .Let us consider problem (1.1)–(1.2) under the following assumptions: •Ais a self-adjoint nonnegative operator on a separable Hilbe rt spaceH, •the coefficient c: [0,+∞)→Rsatisfies the strict hyperbolicity assumption (1.4) and theω-continuity assumption (2.7) for some continuity modulus ω(x), 10•σ∈[0,1/2]andδ>0are two real numbers such that 4δ2µ1>Λ2 ∞+2δΛ∞, (3.4) where we set Λ∞:= limsup ε→0+ω(ε) ε1−2σ, (3.5) •(u0,u1)∈D(A1/2)×H. Letube the unique solution to the problem provided by Theorem A. Thenuactually satisfies u∈C0/parenleftbig [0,+∞),D(A1/2)/parenrightbig ∩C1([0,+∞),H). Moreover, for every ν≥1such that 4δ2µ1≥/bracketleftbigg λ1−2σω/parenleftbigg1 λ/parenrightbigg/bracketrightbigg2 +2δ/bracketleftbigg λ1−2σω/parenleftbigg1 λ/parenrightbigg/bracketrightbigg (3.6) for everyλ≥ν, it turns out that |u′ ν,+(t)|2+2µ1|A1/2uν,+(t)|2≤4|u1,ν,+|2+2(3δ2+µ2)|A1/2u0,ν,+|2(3.7) for everyt≥0. Let us make a few comments on the first two statements. Remark 3.3. Inbothresultsweprovedthatasuitablehigh-frequencycompone nt ofthe solution can be uniformly bounded in terms of initial data. Low-frequ ency components might in general diverge as t→+∞. Nevertheless, they can always be estimated as follows. Let us just assume that c∈L1((0,T)) for every T >0. Then for every ν≥0 the component uν,−(t) satisfies |u′ ν,−(t)|2+|A1/2uν,−(t)|2≤/parenleftbig |u1,ν,−|2+|A1/2u0,ν,−|2/parenrightbig exp/parenleftbigg νt+ν/integraldisplayt 0|c(s)|ds/parenrightbigg (3.8) for everyt≥0. Indeed, let F(t) denote the left-hand side of (3.8). Then F′(t) =−4δ|Aσ/2u′ ν,−(t)|2+2(1−c(t))/a\}⌊ra⌋ketle{tu′ ν,−(t),Auν,−(t)/a\}⌊ra⌋ketri}ht ≤2(1+|c(t)|)·|u′ ν,−(t)|·ν|A1/2uν,−(t)| ≤ν(1+|c(t)|)F(t) for almost every t≥0, so that (3.8) follows by integrating this differential inequality. 11Remark 3.4. The phase spaces involved in Theorem 3.1 and Theorem 3.2 are exactly the same which are known to be optimal when c(t) is constant (see [9]). In particular, the only possible choice in the subcritical regime is the classic energy s paceD(A1/2)×H, or more generally D(Aα+1/2)×D(Aα). This “gap1/2” between the powers of Ainvolved in the phase space is typical of hyperbolic problems, and it is the same which appears in the classic results of section 2. On the contrary, in the supercritical regime there is an interval of possible gaps, described by (1.6). This interval is always centered in 1/2, but also d ifferent values are allowed, including negative ones when σ>1. Remark 3.5. The classic example of application of Theorem 3.2 is the following. Let us assume that c(t) isα-H¨ older continuous for some α∈(0,1), namely ω(x) =Mxαfor some constant M. Then problem (1.1)–(1.2) is well-posed in the energy space provided that either α>1−2σ, orα= 1−2σandMis small enough. Indeed, for α>1−2σwe get Λ∞= 0, and hence (3.4) is automatically satisfied. For α= 1−2σwe get Λ ∞=M, so that (3.4) is satisfied provided that Mis small enough. In all other cases, namely when either α <1−2σ, orα= 1−2σandMis large enough, only Theorem B applies to initial data in Sobolev spaces, prov iding global existence just in the sense of Gevrey ultradistributions of order ( 1−α)−1. Remark 3.6. Let us examine the limit case σ= 0, which falls in the subcritical regime. Whenσ= 0, assumption (3.4) is satisfied if and only if c(t) is Lipschitz continuous and its Lipschitz constant is small enough. On the other hand, in the Lipschitz case it is a classic result that problem (1.1)–(1.2) is well-posed in the energy s pace, regardless of the Lipschitz constant. Therefore, the result stated in Theor em 3.2 is non-optimal whenσ= 0 andc(t) is Lipschitz continuous. A simple refinement of our argument would lead to the full result also in this case, but unfortunately it would be useless in all other limit cases in which c(t) isα-H¨ older continuous with α= 1−2σandσ∈(0,1/2]. We refer to section 4 for further details. Remark 3.7. Let us examine the limit case σ= 1/2, which falls both in the subcritical and in the supercritical regime, so that the conclusions of Theorem 3.1 and Theorem 3.2 coexist. Both of them provide well-posedness in the energy space, but with different assumptions. Theorem 3.1 needs less assumptions on c(t), which is only required to be measurable and to satisfy the degenerate hyperbolicity assumption (1.7), but it requires δto be large enough so that 4 δ2≥µ2. On the contrary, Theorem 3.2 needs less assumptions on δ, which is only required to be positive, but it requires c(t) to be continuous and to satisfy the strict hyperbolicity assumption (1.4). Indeed, inequality (3.4) is automatically satisfied in the caseσ= 1/2 because Λ ∞= 0. The existence of two different sets of assumptions leading to the sa me conclusion suggests the existence of a unifying statement, which could proba bly deserve further investigation. 123.2 Gevrey regularity for positive times A strong dissipation in the range σ∈(0,1) has a regularizing effect on initial data, provided that the solution exists in Sobolev spaces. In the following t wo statements we quantify this effect in terms of scales of Gevrey spaces. Both results can be summed up by saying that the solution lies, for po sitive times, in Gevrey spaces of order (2min {σ,1−σ})−1. It is not difficult to show that this order is optimal, even in the case where c(t) is constant. Theorem 3.8 (Supercritical dissipation) .Let us consider problem (1.1)–(1.2) under the same assumptions of Theorem 3.1, and let ube the unique solution to the problem provided by Theorem A. Let us assume in addition that either σ∈(1/2,1), orσ= 1/2and4δ2>µ2. Let us setϕ(x) :=x2(1−σ), and C(t) :=/integraldisplayt 0c(s)ds. (3.9) Then there exists r>0such that (u,u′)∈C0/parenleftbig (0,+∞),Gϕ,α,rC(t)(A)×Gϕ,β,rC(t)(A)/parenrightbig , (3.10) and there exist ν≥1andK >0such that /⌊ard⌊lu′ ν,+(t)/⌊ard⌊l2 ϕ,β,rC(t)+/⌊ard⌊luν,+(t)/⌊ard⌊l2 ϕ,α,rC(t)≤K/parenleftbig |Aβu1,ν,+|2+|Aαu0,ν,+|2/parenrightbig (3.11) for everyt>0. The constants r,ν, andKdepend only on δ,µ2, andσ. Of course, (3.10) and (3.11) are nontrivial only if C(t)>0, which is equivalent to saying that the coefficient c(t) is not identically 0 in [0 ,t]. On the other hand, this weak form of hyperbolicity is necessary, since no regularizing effect on u(t) can be expected as long asc(t) vanishes. Theorem 3.9 (Subcritical dissipation) .Let us consider problem (1.1)–(1.2) under the same assumptions of Theorem 3.2, and let ube the unique solution to the problem provided by Theorem A. Let us assume in addition that σ∈(0,1/2](instead of σ∈[0,1/2]), and let us set ϕ(x) :=x2σ. Then there exists r>0such that u∈C0/parenleftbig (0,+∞),Gϕ,1/2,rt(A)/parenrightbig ∩C1((0,+∞),Gϕ,0,rt(A)), and there exist ν≥1andK >0such that /⌊ard⌊lu′ ν,+(t)/⌊ard⌊l2 ϕ,0,rt+/⌊ard⌊luν,+(t)/⌊ard⌊l2 ϕ,1/2,rt≤K/parenleftbig |u1,ν,+|2+|A1/2u0,ν,+|2/parenrightbig (3.12) for everyt>0. The constants r,ν, andKdepend only on δ,µ1,µ2,σandω. TheestimateswhichprovideGevreyregularityofhigh-frequencyc omponentsprovide also the decay of the same components as t→+∞. We refer to Lemma 5.1 and Lemma 5.2 for further details. 133.3 Counterexamples The following result shows that even strongly dissipative hyperbolic e quations can ex- hibit the (DGCS)-phenomenon, provided that we are in the subcritic al regime. Theorem 3.10 ((DGCS)-phenomenon) .LetAbe a linear operator on a Hilbert space H. Let us assume that there exists a countable (not necessaril y complete) orthonormal system{ek}inH, and an unbounded sequence {λk}of positive real numbers such that Aek=λ2 kekfor everyk∈N. Letσ∈[0,1/2)andδ>0be real numbers. Letω: [0,+∞)→[0,+∞)be a continuity modulus such that lim ε→0+ω(ε) ε1−2σ= +∞. (3.13) Letϕ: (0,+∞)→(0,+∞)andψ: (0,+∞)→(0,+∞)be two functions such that lim x→+∞x ϕ(x)ω/parenleftbigg1 x/parenrightbigg = lim x→+∞x ψ(x)ω/parenleftbigg1 x/parenrightbigg = +∞. (3.14) Then there exist a function c:R→Rsuch that 1 2≤c(t)≤3 2∀t∈R, (3.15) |c(t)−c(s)| ≤ω(|t−s|)∀(t,s)∈R2, (3.16) and a solution u(t)to equation (1.1) such that (u(0),u′(0))∈ Gϕ,r,1/2(A)×Gϕ,r,0(A)∀r>0, (3.17) (u(t),u′(t))/\e}atio\slash∈ G−ψ,R,1/2(A)×G−ψ,R,0(A)∀R>0,∀t>0.(3.18) Remark 3.11. Due to (3.15), (3.16), and (3.17), the function u(t) provided by Theo- rem 3.10 is a solution to (1.1) in the sense of Theorem A with ψ(x) :=x, or even better in the sense of Theorem B with ψ(x) :=xω(1/x). Remark 3.12. Assumption (3.13) is equivalent to saying that Λ ∞defined by (3.5) is equal to + ∞, so that (3.4) can not be satisfied. In other words, Theorem 3.2 giv es well-posedness in the energy space if Λ ∞is 0 or small, while Theorem 3.10 provides the (DGCS)-phenomenon if Λ ∞= +∞. The case where Λ ∞is finite but large remains open. We suspect that the (DGCS)-phenomenon is still possible, bu t our construction does not work. We comment on this issue in the first part of section 6 . Finally, Theorem 3.10 shows that assumptions (2.9) and (2.10) of The orems B and C are optimal also in the subcritical dissipative case with Λ ∞= +∞. If initial data are in the Gevrey space with ϕ(x) =xω(1/x), solutions remain in the same space. If initial are in a Gevrey space corresponding to some ϕ(x)≪xω(1/x), then it may happen that for positive times the solution lies in the space of ultradistributions with ψ(x) :=xω(1/x), but not in the space of ultradistributions corresponding to any give nψ(x)≪xω(1/x). 144 Heuristics Thefollowingpicturessummarizeroughlytheresultsofthispaper. I nthehorizontalaxis we represent the time-regularity of c(t). With some abuse of notation, values α∈(0,1) mean that c(t) isα-H¨ older continuous, α= 1 means that it is Lipschitz continuous, α >1 means even more regular. In the vertical axis we represent the s pace-regularity of initial data, where the value sstands for the Gevrey space of order s(so that higher values ofsmean lower regularity). The curve is s= (1−α)−1. α 1s 1 δ= 0Potential (DGCS)-phenomenon Well-posedness α 1−2σ/Bullets 1 δ >0,0<σ<1/2α 1s 1 δ >0, σ>1/2 Forδ= 0 we have the situation described in Remark 2.5 and Remark 2.6, name ly well-posedness provided that either c(t) is Lipschitz continuous or c(t) isα-H¨ older con- tinuous and initial data are in Gevrey spaces of order less than or eq ual to (1 −α)−1, and (DGCS)-phenomenon otherwise. The same picture applies if δ >0 andσ= 0. Whenδ >0 and 0< σ <1/2, the full strip with α >1−2σfalls in the well- posedness region, as stated in Theorem 3.2. The region with α <1−2σis divided as in the non-dissipative case. Indeed, Theorem C still provides well-po sedness below the curve and on the curve, while Theorem 3.10 provides the (DGCS)-ph enomenon above the curve. What happens on the vertical half-line which separates the two regions is less clear (it is the region where Λ ∞is positive and finite, see Remark 3.12). Finally, when δ >0 andσ>1/2 well-posedness dominates because of Theorem 3.1, even in the degenerate hyperbolic case. Now we present a rough justification of this threshold effect. As alr eady observed, existence results for problem (1.1)–(1.2) are related to estimates for solutions to the family of ordinary differential equations (1.8). Let us consider the simplest energy function E(t) :=|u′ λ(t)|2+λ2|uλ(t)|2, whose time-derivative is E′(t) =−4δλ2σ|u′ λ(t)|2+2λ2(1−c(t))uλ(t)u′ λ(t) ≤ −4δλ2σ|u′ λ(t)|2+λ(1+|c(t)|)E(t). (4.1) 15Sinceδ≥0, a simple integration gives that E(t)≤ E(0)exp/parenleftbigg λt+λ/integraldisplayt 0|c(s)|ds/parenrightbigg , (4.2) which is almost enough to establish Theorem A. Ifinaddition c(t)isω-continuousandsatisfiesthestricthyperbolicitycondition(1.4), then (4.2) can be improved to E(t)≤M1E(0)exp(M2λω(1/λ)t) (4.3) for suitable constants M1andM2. Estimates of this kind are the key point in the proof of both Theorem B and Theorem C. Moreover, the (DGCS)-phenom enon is equivalent to saying that the term λω(1/λ) in (4.3) is optimal. Let us assume now that δ >0. Ifσ >1/2, orσ= 1/2 andδis large enough, then it is reasonable to expect that the first (negative) term in the right-hand side of (4.1) dominates the second one, and hence E(t)≤ E(0), which is enough to establish well-posedness in Sobolev spaces. Theorem 3.1 confirms this intuition . Ifσ≤1/2 andc(t) is constant, then (1.8) can be explicitly integrated, obtaining that E(t)≤ E(0)exp/parenleftbig −2δλ2σt/parenrightbig . (4.4) Ifc(t) isω-continuous and satisfies the strict hyperbolicity assumption (1.4) , then we expect a superposition of the effects of the coefficient, repres ented by (4.3), and the effects of the damping, represented by (4.4). We end up with E(t)≤M1E(0)exp/parenleftbig [M2λω(1/λ)−2δλ2σ]t/parenrightbig . (4.5) Therefore, it is reasonable to expect that E(t) satisfies an estimate independent of λ, which guarantees well-posedness in Sobolev spaces, provided tha tλω(1/λ)≪λ2σ, or λω(1/λ)∼λ2σandδis large enough. Theorem 3.2 confirms this intuition. The same argument applies if σ= 0 andω(x) =Lx, independently of L(see Remark 3.6). On the contrary, in all other cases the right-hand side of (4.5) dive rges asλ→ +∞, opening the door to the (DGCS)-phenomenon. We are able to show that it does happen provided that λω(1/λ)≫λ2σ. We refer to the first part of section 6 for further comments. 5 Proofs of well-posedness and regularity results All proofs of our main results concerning well-posedness and regula rity rely on suitable estimates for solutions to the ordinary differential equation (1.8) w ith initial data uλ(0) =u0, u′ λ(0) =u1. (5.1) For the sake of simplicity in the sequel we write u(t) instead of uλ(t). 165.1 Supercritical dissipation Let us consider the case σ≥1/2. The key tool is the following. Lemma 5.1. Let us consider problem (1.8)–(5.1) under the following ass umptions: •the coefficient c: [0,+∞)→Ris measurable and satisfies the degenerate hyper- bolicity assumption (1.7), •δ,λ,σare positive real numbers such that 4δ2λ4σ−2≥µ2. (5.2) Then the solution u(t)satisfies the following estimates. (1) For every t≥0it turns out that |u(t)|2≤2 δ2λ4σu2 1+3u2 0, (5.3) |u′(t)|2≤/parenleftbigg 2+µ2 2 δ4λ8σ−4/parenrightbigg u2 1+3µ2 2 2δ2λ4σ−4u2 0. (5.4) (2) Let us assume in addition that λ≥1andσ≥1/2, and letαandβbe two real numbers satisfying (1.6). Then for every t≥0it turns out that λ4β|u′(t)|2+λ4α|u(t)|2≤/parenleftbigg 2+2 δ2+µ2 2 δ4/parenrightbigg λ4βu2 1+3/parenleftbigg 1+µ2 2 2δ2/parenrightbigg λ4αu2 0.(5.5) (3) In addition to the assumptions of the statement (2), let u s assume also that there existsr>0satisfying the following three inequalities: δλ4σ−2>rµ2,2δr≤1,4δ2λ4σ−2≥(1+2rδ)µ2.(5.6) Then for every t≥0it turns out that λ4β|u′(t)|2+λ4α|u(t)|2≤/bracketleftbigg 2/parenleftbigg 1+2µ2 2 δ4+1 δ2/parenrightbigg λ4βu2 1+3/parenleftbigg 1+2µ2 2 δ2/parenrightbigg λ4αu2 0/bracketrightbigg × ×exp/parenleftbigg −2rλ2(1−σ)/integraldisplayt 0c(s)ds/parenrightbigg . (5.7) 17ProofLet us consider the energy E(t) defined in (1.9). Since −3 4|u′(t)|2−4 3δ2λ4σ|u(t)|2≤2δλ2σu(t)u′(t)≤ |u′(t)|2+δ2λ4σ|u(t)|2, we easily deduce that 1 4|u′(t)|2+2 3δ2λ4σ|u(t)|2≤E(t)≤2|u′(t)|2+3δ2λ4σ|u(t)|2∀t≥0.(5.8) Statement (1) The time-derivative of E(t) is E′(t) =−2/parenleftbig δλ2σ|u′(t)|2+δλ2σ+2c(t)|u(t)|2+λ2c(t)u(t)u′(t)/parenrightbig .(5.9) The right-hand side is a quadratic form in u(t) andu′(t). The coefficient of |u′(t)|2 is negative. Therefore, this quadratic form is less than or equal to 0 for all values of u(t) andu′(t) if and only if 4δ2λ4σ−2c(t)≥c2(t), and this is always true because of (1.7) and (5.2). It follows that E′(t)≤0 for (almost) everyt≥0, and hence δ2λ4σ|u(t)|2≤E(t)≤E(0)≤2u2 1+3δ2λ4σu2 0, (5.10) which is equivalent to (5.3). In order to estimate u′(t), we rewrite (1.8) in the form u′′(t)+2δλ2σu′(t) =−λ2c(t)u(t), which we interpret as a first order linear equation with constant coe fficients in the unknownu′(t), with the right-hand side as a forcing term. Integrating this differ ential equation in u′(t), we obtain that u′(t) =u1exp/parenleftbig −2δλ2σt/parenrightbig −/integraldisplayt 0λ2c(s)u(s)exp/parenleftbig −2δλ2σ(t−s)/parenrightbig ds. (5.11) From (1.7) and (5.3) it follows that |u′(t)| ≤ |u1|+µ2λ2·max t∈[0,T]|u(t)|·/integraldisplayt 0e−2δλ2σ(t−s)ds ≤ |u1|+µ2λ2 2δλ2σ/parenleftbigg2 δ2λ4σu2 1+3u2 0/parenrightbigg1/2 , and therefore |u′(t)|2≤2|u1|2+µ2 2λ4 2δ2λ4σ/parenleftbigg2 δ2λ4σu2 1+3u2 0/parenrightbigg , which is equivalent to (5.4). 18Statement (2) Exploiting (5.3) and (5.4), with some simple algebra we obtain that λ4β|u′(t)|2+λ4α|u(t)|2≤/parenleftbigg 2+µ2 2 δ4·1 λ4(2σ−1)+2 δ2·1 λ4(β+σ−α)/parenrightbigg λ4βu2 1 +3/parenleftbigg 1+µ2 2 2δ2·1 λ4(α−β+σ−1)/parenrightbigg λ4αu2 0. All exponents of λ’s in denominators are nonnegative owing to (1.6). Therefore, sinceλ≥1, all those fractions can be estimated with 1. This leads to (5.5). Statement (3) Let us define C(t) as in (3.9). To begin with, we prove that in this case the function E(t) satisfies the stronger differential inequality E′(t)≤ −2rλ2(1−σ)c(t)E(t), (5.12) and hence E(t)≤E(0)exp/parenleftbig −2rλ2(1−σ)C(t)/parenrightbig ∀t≥0. (5.13) Coming back to (5.9), inequality (5.12) is equivalent to λ2σ/parenleftbig δ−rλ2−4σc(t)/parenrightbig |u′(t)|2+δλ2σ+2(1−2rδ)c(t)|u(t)|2+λ2(1−2rδ)c(t)u(t)u′(t)≥0. As in the proof of statement (1), we consider the whole left-hand s ide as a quadratic form inu(t) andu′(t). Sincec(t)≤µ2, from the first inequality in (5.6) it follows that δλ4σ−2>rµ2≥rc(t), which is equivalent to saying that the coefficient of |u′(t)|2is positive. Therefore, the quadratic form is nonnegative for all values of u(t) andu′(t) if and only if 4λ2σ/parenleftbig δ−rλ2−4σc(t)/parenrightbig ·δλ2σ+2c(t)(1−2rδ)≥λ4c2(t)(1−2rδ)2, hence if and only if (1−2rδ)c(t)/bracketleftbig 4δ2λ4σ−2−(1+2rδ)c(t)/bracketrightbig ≥0, and this follows from (1.7) and from the last two inequalities in (5.6). Now from (5.13) it follows that δ2λ4σ|u(t)|2≤E(t)≤E(0)exp/parenleftbig −2rλ2(1−σ)C(t)/parenrightbig , (5.14) which provides an estimate for |u(t)|. In order to estimate u′(t), we write it in the form (5.11), and we estimate the two terms separately. The third inequa lity in (5.6) implies that 2δλ4σ−2≥rµ2. SinceC(t)≤µ2t, it follows that 2δλ2σt≥rλ2−2σµ2t≥rλ2−2σC(t), 19and hence /vextendsingle/vextendsingleu1exp/parenleftbig −2δλ2σt/parenrightbig/vextendsingle/vextendsingle≤ |u1|exp/parenleftbig −2δλ2σt/parenrightbig ≤ |u1|exp/parenleftbig −rλ2(1−σ)C(t)/parenrightbig .(5.15) As for the second terms in (5.11), we exploit (5.14) and we obtain tha t /vextendsingle/vextendsingle/vextendsingle/vextendsingle/integraldisplayt 0λ2c(s)u(s)exp/parenleftbig −2δλ2σ(t−s)/parenrightbig ds/vextendsingle/vextendsingle/vextendsingle/vextendsingle≤λ2µ2/integraldisplayt 0|u(s)|exp/parenleftbig −2δλ2σ(t−s)/parenrightbig ds ≤µ2[E(0)]1/2 δλ2σ−2exp/parenleftbig −2δλ2σt/parenrightbig/integraldisplayt 0exp/parenleftbig −rλ2(1−σ)C(s)+2δλ2σs/parenrightbig ds. From the first inequality in (5.6) it follows that 2δλ2σ−rλ2(1−σ)c(s)≥2δλ2σ−rλ2(1−σ)µ2≥δλ2σ, hence /integraldisplayt 0exp/parenleftbig −rλ2(1−σ)C(s)+2δλ2σs/parenrightbig ds ≤1 δλ2σ/integraldisplayt 0/parenleftbig 2δλ2σ−rλ2(1−σ)c(s)/parenrightbig exp/parenleftbig 2δλ2σs−rλ2(1−σ)C(s)/parenrightbig ds ≤1 δλ2σexp/parenleftbig 2δλ2σt−rλ2(1−σ)C(t)/parenrightbig , and therefore /vextendsingle/vextendsingle/vextendsingle/vextendsingle/integraldisplayt 0λ2c(s)u(s)exp/parenleftbig −2δλ2σ(t−s)/parenrightbig ds/vextendsingle/vextendsingle/vextendsingle/vextendsingle≤µ2[E(0)]1/2 δ2λ4σ−2exp/parenleftbig −rλ2(1−σ)C(t)/parenrightbig .(5.16) From (5.11), (5.15) and (5.16) we deduce that |u′(t)| ≤/parenleftbigg |u1|+µ2[E(0)]1/2 δ2λ4σ−2/parenrightbigg exp/parenleftbig −rλ2(1−σ)C(t)/parenrightbig , and hence |u′(t)|2≤/parenleftbigg 2|u1|2+2µ2 2E(0) δ4λ8σ−4/parenrightbigg exp/parenleftbig −2rλ2(1−σ)C(t)/parenrightbig . (5.17) Finally, we estimate E(0) as in (5.10). At this point, estimate (5.7) follows from (5.17) and (5.14) with some simple algebra (we need to exploit that λ≥1 and assump- tion (1.6) exactly as in the proof of statement (2)). /square 205.1.1 Proof of Theorem 3.1 Let us fix a real number ν≥1 such that 4 δ2ν4σ−2≥µ2(such a number exists because of ourassumptions on δandσ). Letusconsiderthecomponents uk(t)ofu(t)corresponding to eigenvalues λk≥ν. Sinceλk≥1 and 4δ2λ4σ−2 k≥µ2, we can apply statement (2) of Lemma 5.1 to these components. If u0kandu1kdenote the corresponding components of initial data, estimate (5.5) read as λ4β k|u′ k(t)|2+λ4α k|uk(t)|2≤/parenleftbigg 2+2 δ2+µ2 2 δ4/parenrightbigg λ4β k|u1,k|2+3/parenleftbigg 1+µ2 2 2δ2/parenrightbigg λ4α k|u0,k|2. Summing over all λk≥νwe obtain exactly (3.3). This proves that uν,+(t) is bounded with values in D(Aα) andu′ ν,+(t) is bounded with values in D(Aβ). The same estimate guarantees the uniform convergence in the whole half-line t≥0 of the series defining Aαuν,+(t) andAβu′ ν,+(t). Since all summands are continuous, and the convergence is uniform, the sum is continu ous as well. Since low-frequency components uν,−(t) andu′ ν,−(t) are continuous (see Remark 3.3), (3.2) is proved. /square 5.1.2 Proof of Theorem 3.8 Let us fix a real number ν≥1 such that 4 δ2ν4σ−2>µ2(such a number exists because of our assumptions on δandσ). Then there exists r>0 such that the three inequalities in (5.6) hold true for every λ≥ν. Therefore, we can apply statement (3) of Lemma 5.1 to all components uk(t) ofu(t) corresponding to eigenvalues λk≥ν. Ifu0kandu1kdenote the corresponding components of initial data, estimate (5.7) read as /parenleftBig λ4β k|u′ k(t)|2+λ4α k|uk(t)|2/parenrightBig exp/parenleftbigg 2rλ2(1−σ) k/integraldisplayt 0c(s)ds/parenrightbigg ≤K/parenleftBig λ4β k|u1k|2+λ4α k|u0k|2/parenrightBig for everyt≥0, whereKis a suitable constant depending only on µ2andδ. Summing over allλk≥νwe obtain exactly (3.11). The continuity of u(t) andu′(t) with values in the suitable spaces follows from the uniform convergence of serie s as in the proof of Theorem 3.1. /square 5.2 Subcritical dissipation Let us consider the case 0 ≤σ≤1/2. The key tool is the following. Lemma 5.2. Let us consider problem (1.8)–(5.1) under the following ass umptions: •the coefficient c: [0,+∞)→Rsatisfies the strict hyperbolicity assumption (1.4) and theω-continuity assumption (2.7) for some continuity modulus ω(x), •δ>0,λ>0, andσ≥0are real numbers satisfying (3.6). 21Then the solution u(t)satisfies the following estimates. (1) It turns out that |u′(t)|2+2λ2µ1|u(t)|2≤4u2 1+2/parenleftbig 3δ2λ4σ+λ2µ2/parenrightbig u2 0∀t≥0.(5.18) (2) Let us assume in addition that λ≥1,σ∈[0,1/2], and there exists a constant r∈(0,δ)such that 4(δ−r)(δµ1−rµ2)≥/bracketleftbigg λ1−2σω/parenleftbigg1 λ/parenrightbigg/bracketrightbigg2 +2δ(1+2r)/bracketleftbigg λ1−2σω/parenleftbigg1 λ/parenrightbigg/bracketrightbigg +8rδ3.(5.19) Then for every t≥0it turns out that |u′(t)|2+2λ2µ1|u(t)|2≤/bracketleftbig 4u2 1+2/parenleftbig 3δ2λ4σ+λ2µ2/parenrightbig u2 0/bracketrightbig exp/parenleftbig −2rλ2σt/parenrightbig .(5.20) ProofFor everyε>0 we introduce the regularized coefficient cε(t) :=1 ε/integraldisplayt+ε tc(s)ds∀t≥0. It is easy to see that cε∈C1([0,+∞),R) and satisfies the following estimates: µ1≤cε(t)≤µ2∀t≥0, (5.21) |c(t)−cε(t)| ≤ω(ε)∀t≥0, (5.22) |c′ ε(t)| ≤ω(ε) ε∀t≥0. (5.23) Approximated energy For everyε >0 we consider the approximated hyperbolic energyEε(t) defined in (1.10). Since −1 2|u′(t)|2−2δ2λ4σ|u(t)|2≤2δλ2σu(t)u′(t)≤ |u′(t)|2+δ2λ4σ|u(t)|2, we deduce that 1 2|u′(t)|2+µ1λ2|u(t)|2≤Eε(t)≤2|u′(t)|2+(3δ2λ4σ+λ2µ2)|u(t)|2(5.24) for everyε>0 andt≥0. The time-derivative of Eε(t) is E′ ε(t) =−2δλ2σ|u′(t)|2−2δλ2σ+2c(t)|u(t)|2 −2λ2(c(t)−cε(t))u(t)u′(t)+λ2c′ ε(t)|u(t)|2, (5.25) hence E′ ε(t)≤ −2δλ2σ|u′(t)|2−/parenleftbig 2δλ2σ+2c(t)−λ2|c′ ε(t)|/parenrightbig |u(t)|2 +2λ2|c(t)−cε(t)|·|u(t)|·|u′(t)|. (5.26) 22Statement (1) We claim that, for a suitable choice of ε, it turns out that E′ ε(t)≤0∀t≥0. (5.27) If we prove this claim, then we apply (5.24) with that particular value o fεand we obtain that 1 2|u′(t)|2+µ1λ2|u(t)|2≤Eε(t)≤Eε(0)≤2u2 1+(3δ2λ4σ+λ2µ2)u2 0, which is equivalent to (5.18). Inordertoprove(5.27),weconsiderthewholeright-handsideof( 5.26)asaquadratic form in|u(t)|and|u′(t)|. Since the coefficient of |u′(t)|2is negative, this quadratic form is less than or equal to 0 for all values of |u(t)|and|u′(t)|if and only if 2δλ2σ·/parenleftbig 2δλ2σ+2c(t)−λ2|c′ ε(t)|/parenrightbig −λ4|c(t)−cε(t)|2≥0, hence if and only if 4δ2λ4σ−2c(t)≥ |c(t)−cε(t)|2+2δλ2σ−2|c′ ε(t)|. (5.28) Now in the left-hand side we estimate c(t) from below with µ1, and we estimate from above the terms in the right-hand side as in (5.22) and (5.23). We obt ain that (5.28) holds true whenever 4δ2µ1≥ω2(ε) λ4σ−2+2δω(ε) λ2σε. This condition is true when ε:= 1/λthanks to assumption (3.6). This completes the proof of (5.18). Statement (2) Let us assume now that λ≥1 and that (5.19) holds true for some r∈(0,δ). In this case we claim that, for a suitable choice of ε>0, the stronger estimate E′ ε(t)≤ −2rλ2σEε(t)∀t≥0 (5.29) holds true, hence Eε(t)≤Eε(0)exp/parenleftbig −2rλ2σt/parenrightbig ∀t≥0. Due to (5.24), this is enough to deduce (5.20). So it remains to prove (5.29). Owing to (5.25), inequality (5.29) is equivalent to 2λ2σ(δ−r)|u′(t)|2+/bracketleftbig 2λ2σ+2(δc(t)−rcε(t))−λ2c′ ε(t)−4rδ2λ6σ/bracketrightbig |u(t)|2 +2/bracketleftbig λ2(c(t)−cε(t))−2rδλ4σ/bracketrightbig u(t)u′(t)≥0. Keeping (1.4) and (5.21) into account, the last inequality is proved if w e show that 2λ2σ(δ−r)|u′(t)|2+/bracketleftbig 2λ2σ+2(δµ1−rµ2)−λ2|c′ ε(t)|−4rδ2λ6σ/bracketrightbig |u(t)|2 23−2/bracketleftbig λ2|c(t)−cε(t)|+2rδλ4σ/bracketrightbig |u(t)|·|u′(t)| ≥0. As in the proof of the first statement, we consider the whole left-h and side as a quadratic form in |u(t)|and|u′(t)|. The coefficient of |u′(t)|is positive because r < δ. Therefore, this quadratic form is nonnegative for all values of |u(t)|and|u′(t)|if and only if 2λ2σ(δ−r)·/bracketleftbig 2λ2σ+2(δµ1−rµ2)−λ2|c′ ε(t)|−4rδ2λ6σ/bracketrightbig ≥/bracketleftbig λ2|c(t)−cε(t)|+2rδλ4σ/bracketrightbig2. Now we rearrange the terms, and we exploit (5.22) and (5.23). We ob tain that the last inequality is proved if we show that 4(δ−r)(δµ1−rµ2)≥λ2−4σω2(ε)+2δ/parenleftbig 1+2rελ2σ/parenrightbigω(ε) ελ2σ+8rδ3 λ2−4σ.(5.30) Finally, we choose ε:= 1/λ, so that (5.30) becomes 4(δ−r)(δµ1−rµ2)≥/bracketleftbigg λ1−2σω/parenleftbigg1 λ/parenrightbigg/bracketrightbigg2 +2δ/parenleftbigg 1+2r λ1−2σ/parenrightbigg/bracketleftbigg λ1−2σω/parenleftbigg1 λ/parenrightbigg/bracketrightbigg +8rδ3 λ2−4σ. Sinceλ≥1 andσ≤1/2, this inequality follows from assumption (5.19). /square 5.2.1 Proof of Theorem 3.2 Let us rewrite (3.5) in the form Λ∞= limsup λ→+∞λ1−2σω/parenleftbigg1 λ/parenrightbigg . (5.31) Due to (3.4), there exists ν≥1 such that (3.6) holds true for every λ≥ν. Therefore, we can apply statement (1) of Lemma 5.2 to the components uk(t) ofu(t) corresponding to eigenvalues λk≥ν. Ifu0kandu1kdenote the corresponding components of initial data, estimate (5.18) read as |u′ k(t)|2+2λ2 kµ1|uk(t)|2≤4|u1k|2+2/parenleftbig 3δ2λ4σ k+λ2 kµ2/parenrightbig |u0k|2. Sinceσ≤1/2 and we chose ν≥1, this implies that |u′ k(t)|2+2λ2 kµ1|uk(t)|2≤4|u1k|2+2/parenleftbig 3δ2+µ2/parenrightbig λ2 k|u0k|2. Summing over all λk≥νwe obtain exactly (3.7). This proves that uν,+(t) is bounded with values in D(A1/2) andu′ ν,+(t) is bounded with values in H. The continuity of u(t) andu′(t) with values in the same spaces follows from the uniform convergence of series as in the proof of Theorem 3.1./square 245.2.2 Proof of Theorem 3.9 Let us rewrite (3.5) in the form (5.31). Due to (3.4), there exists r >0 andν≥1 such that (5.19) holds true for every λ≥ν. Therefore, we can apply statement (2) of Lemma 5.2 to the components uk(t) ofu(t) corresponding to eigenvalues λk≥ν. Ifu0k andu1kdenote the corresponding components of initial data, estimate (5 .20) reads as /parenleftbig |u′ k(t)|2+2λ2 kµ1|uk(t)|2/parenrightbig exp/parenleftbig 2rλ2σ kt/parenrightbig ≤4|u1k|2+2/parenleftbig 3δ2λ4σ k+λ2 kµ2/parenrightbig |u0k|2. Sinceσ≤1/2 and we chose ν≥1, this implies that /parenleftbig |u′ k(t)|2+2λ2 kµ1|uk(t)|2/parenrightbig exp/parenleftbig 2rλ2σ kt/parenrightbig ≤4|u1k|2+2/parenleftbig 3δ2+µ2/parenrightbig λ2 k|u0k|2 for everyt≥0. Summing over all λk≥νwe obtain (3.12) with a constant Kdepending only onµ1,µ2, andδ. The continuity of u(t) andu′(t) with values in the suitable spaces follows from the uniform convergence of series as in the proof of Th eorem 3.1. /square 6 The (DGCS)-phenomenon In this section we prove Theorem 3.10. Let us describe the strateg y before entering into details. Roughly speaking, what we need is a solution u(t) whose components uk(t) are small at time t= 0 and huge at time t>0. The starting point is given by the following functions b(ε,λ,t) := (2ελ−δλ2σ)t−εsin(2λt), w(ε,λ,t) := sin(λt)exp(b(ε,λ,t)), (6.1) γ(ε,λ,t) := 1+δ2 λ2−4σ−16ε2sin4(λt)−8εsin(2λt). (6.2) With some computations it turns out that w′′(ε,λ,t)+2δλ2σw′(ε,λ,t)+λ2γ(ε,λ,t)w(ε,λ,t) = 0 ∀t∈R, where “primes” denote differentiation with respect to t. As a consequence, if we set c(t) :=γ(ε,λ,t) andε:=ω(1/λ), the function u(t) :=w(ε,λ,t) turns out to be a solution to (1.8) which grows as the right-hand side of (4.5). Unfort unately this is not enough, because we need to realize a similar growth for countably ma ny components with the same coefficient c(t). To this end, we argue as in [6]. We introduce a suitable decreasing sequ encetk→0+, and in the interval [ tk,tk−1] we design the coefficient c(t) so thatuk(tk) is small and uk(tk−1) is huge. Then we check that the piecewise defined coefficient c(t) has the required time-regularity, and that uk(t) remains small for t∈[0,tk] and remains huge fort≥tk−1. This completes the proof. Roughly speaking, the coefficient c(t) plays on infinitely many time-scales in order to “activate” countably many components, but these countably m any actions take place 25onebyoneindisjointtimeintervals. Ofcoursethismeansthatthelen gthstk−1−tkofthe “activationintervals”tendto0as k→+∞. Inordertoobtainenoughgrowth, despiteof the vanishing length of activationintervals, we areforced to assum e thatλω(1/λ)≫λ2σ asλ→+∞. In addition, components do not grow exactly as exp( λω(1/λ)t), but just more than exp( ϕ(λ)t) and exp(ψ(λ)t). This is the reason why this strategy does not work when λω(1/λ)∼λ2σandδ is small. In this case one would need components growing exactly as ex p(λω(1/λ)t), but this requires activation intervals of non-vanishing length, which are thus forced to overlap. In a certain sense, the coefficient c(t) should work once againoninfinitely many time-scales, but now the countably many actions should take place in the same time. Definition of sequences From (3.13) and (3.14) it follows that lim x→+∞x1−2σω/parenleftbigg1 x/parenrightbigg = +∞, (6.3) lim x→+∞1 x1−2σω(1/x)+ϕ(x) xω(1/x)+ψ(x) xω(1/x)= 0, (6.4) and a fortiori lim x→+∞x1+2σω/parenleftbigg1 x/parenrightbigg = +∞, (6.5) lim x→+∞x2σ+ϕ(x)+ψ(x) xω/parenleftbigg1 x/parenrightbigg = 0. (6.6) Let us consider the sequence {λk}, which we assumed to be unbounded. Due to (6.5) and (6.4) we can assume, up to passing to a subsequence (not relabeled), that the following inequalities hold true for every k≥1: λk>4λk−1, (6.7) λ1+2σ kω/parenleftbigg1 λk/parenrightbigg ≥δ4 210π21 λ2−8σ k−1+4k2 π2λ2 k−1, (6.8) λ1+2σ kω/parenleftbigg1 λk/parenrightbigg ≥4k2 π2λ3 k−1/parenleftbig λ2σ k−1+ϕ(λk−1)+ψ(λk−1)/parenrightbig ω/parenleftbigg1 λk−1/parenrightbigg ,(6.9) λ1+2σ kω/parenleftbigg1 λk/parenrightbigg ≥λk−1/parenleftbig λ2σ k−1+ϕ(λk−1)+ψ(λk−1)/parenrightbig ω/parenleftbigg1 λk−1/parenrightbigg ,(6.10) 1 λ1−2σ kω(1/λk)+ϕ(λk) λkω(1/λk)+ψ(λk) λkω(1/λk)≤π2 4k21 λ2 k−1. (6.11) Now let us set tk:=4π λk, s k:=π λk/floorleftbigg 2λk λk−1/floorrightbigg , (6.12) 26where⌊α⌋denotes the largest integer less than or equal to α, and εk:=/braceleftbiggλ2σ k+ϕ(λk)+ψ(λk) λkω/parenleftbigg1 λk/parenrightbigg/bracerightbigg1/2 . Properties of the sequences We collect in this section of the proof all the properties of the sequences which are needed in the sequel. First of all, it is clear thatλk→+∞, hencetk→0 andεk→0 (because of (6.6)). Moreover it turns out that tk−1 4=π λk−1≤sk≤2π λk−1=tk−1 2. (6.13) Keeping (6.7) into account, it follows that tk<sk<tk−1∀k≥1, and in particular also sk→0. In addition, it turns out that sin(λktk) = sin(λksk) = 0 (6.14) and |cos(λktk)|=|cos(λksk)|= 1. (6.15) Sinceσ <1/2,λk→+∞,εk→0,tk→0, keeping (6.3) and (6.4) into account, we deduce that the following seven inequalities are satisfied provided th atkis large enough: δ2 λ2−4σ k+16ε2 k+8εk≤1 2, (6.16) εk≤1 4, (6.17) 16πεk+16πδ λ1−2σ k≤2π, (6.18) 1 λ1−2σ kω(1/λk)+ϕ(λk) λkω(1/λk)+ψ(λk) λkω(1/λk)≤1 52·210·π2, (6.19) δ2 (4π)2−4σ(2tk)1−2σsup/braceleftbiggx1−2σ ω(x):x∈(0,tk)/bracerightbigg ≤1 5, (6.20) λ1−2σ kω/parenleftbigg1 λk/parenrightbigg ≥δ2, (6.21) 2δ2 λ2−4σ k−1ω(1/λk−1)≤1 5. (6.22) 27Letk0∈Nbe a positive integer such that (6.16) through (6.22) hold true for e very k≥k0. From (6.21) it follows that εkλk≥δλ2σ k∀k≥k0. (6.23) From (6.19) it follows that 32πεk ω(1/λk)≤1 5∀k≥k0. (6.24) Sincesk≥π/λk−1(see the estimate from below in (6.13)), from (6.8) it follows that εkλksk≥δ2 321 λ2−4σ k−1∀k≥k0, (6.25) εkλksk≥2k∀k≥k0, (6.26) from (6.9) it follows that εkλksk≥2kεk−1λk−1∀k≥k0, (6.27) and from (6.11) it follows that εkλksk≥2k/parenleftbig λ2σ k+ϕ(λk)+ψ(λk)/parenrightbig ∀k≥k0. (6.28) As a consequence of (6.26) through (6.28) it turns out that 2εkλksk≥kεk−1λk−1+2k/parenleftbig λ2σ k+ϕ(λk)+ψ(λk)/parenrightbig +k∀k≥k0.(6.29) Finally, from (6.10) it follows that εkλk≥εk−1λk−1∀k≥k0. (6.30) Definition of c(t)andu(t) For every k≥1, letℓk:R→Rbe defined by ℓk(t) :=δ2 tk−1−sk/parenleftbigg1 λ2−4σ k−1−1 λ2−4σ k/parenrightbigg (t−sk)+1+δ2 λ2−4σ k∀t∈R. Thanks to (6.14), ℓk(t) is the affine function such that ℓk(sk) =γ(εk,λk,sk) and ℓk(tk−1) =γ(εk−1,λk−1,tk−1). Letk0∈Nbe such that (6.16) through (6.22) hold true for every k≥k0. Let us set c(t) := 1 if t≤0, γ(εk,λk,t) ift∈[tk,sk] for somek≥k0, ℓk(t) if t∈[sk,tk−1] for somek≥k0+1, γ(εk0,λk0,sk0) ift≥sk0. 28The following picture describes this definition. The coefficient c(t) is constant for t≤0 and fort≥sk0. In the intervals [ tk,sk] it coincides with γ(εk,λk,t), hence it oscillates, with period of order λ−1 kand amplitude of order εk, around a value which tends to 1. In the intervals [ sk,tk−1] it is just the affine interpolation of the values at the endpoints. /Bullet /Bullet /Bullet /Bullet /Bullet /Bullet /Bullet sk+2tk+1sk+1tksk tk−1sk−1period∼λ−1 kperiod∼λ−1 k−1 1+δ2 λ2−4σ k∼εk∼εk−1 For everyk≥k0, let us consider the solution uk(t) to the Cauchy problem u′′ k(t)+2δλ2σ ku′ k(t)+λ2 kc(t)uk(t) = 0, with “initial” data uk(tk) = 0, u′ k(tk) =λkexp/parenleftbig (2εkλk−δλ2σ k)tk/parenrightbig . (6.31) Then we set ak:=1 kλkexp(−kϕ(λk)), (6.32) and finally u(t) :=∞/summationdisplay k=k0akuk(t)ek. Weclaim that c(t) satisfies (3.15)and(3.16), andthat u(t)satisfies (3.17)and(3.18). The rest of the proof is a verification of these claims. Estimate and continuity of c(t) We prove that |c(t)−1| ≤1 2∀t≥0, (6.33) 29which is equivalent to (3.15), and that c(t) is continuous on the whole real line. To this end, it is enough to check (6.33) in the intervals [ tk,sk], because in the intervals [sk,tk−1]the function c(t) isjust aninterpolationofthevalues at theendpoints, and it is constant for t≤0 and fort≥sk0. In the intervals [ tk,sk] the function c(t) coincides with γ(εk,λk,t), hence from (6.2) it turns out that |c(t)−1|=|γ(εk,λk,t)−1| ≤δ2 λ2−4σ k+16ε2 k+8εk, (6.34) so that (6.33) follows immediately from (6.16). Since the right-hand side of (6.34) tends to 0 as k→+∞, the same estimate shows also thatc(t)→1 ast→0+, which proves the continuity of c(t) int= 0, the only point in which continuity was nontrivial. Estimate on c′(t) We prove that |c′(t)| ≤32εkλk∀t∈(tk,sk),∀k≥k0, (6.35) |c′(t)| ≤32εkλk∀t∈(sk,tk−1),∀k≥k0+1. (6.36) Indeed in the interval ( tk,sk) it turns out that |c′(t)|=|γ′(εk,λk,t)|=/vextendsingle/vextendsingle−64ε2 kλksin3(λkt)cos(λkt)−16εkλkcos(2λkt)/vextendsingle/vextendsingle ≤64ε2 kλk+16εkλk= 16εkλk(4εk+1), so that (6.35) follows from (6.17). In the interval ( sk,tk−1) it turns out that |c′(t)|=δ2 tk−1−sk/parenleftbigg1 λ2−4σ k−1−1 λ2−4σ k/parenrightbigg ≤δ2 tk−1−sk·1 λ2−4σ k−1≤δ2 sk·1 λ2−4σ k−1, where the last inequality follows from the estimate from above in (6.13 ). At this point (6.36) is equivalent to (6.25). Modulus of continuity of c(t) Let us prove that c(t) satisfies (3.16). Since c(t) is continuous, and constant for t≤0 andt≥sk0, it is enough to verify its ω-continuity in (0,sk0]. In turn, the ω-continuity in (0 ,sk0] is proved if we show that |c(ti)−c(tj)| ≤1 5ω(|ti−tj|)∀i≥k0,∀j≥k0, (6.37) |c(a)−c(b)| ≤1 5ω(|a−b|)∀(a,b)∈[tk,sk]2,∀k≥k0, (6.38) 30|c(a)−c(b)| ≤1 5ω(|a−b|)∀(a,b)∈[sk,tk−1]2,∀k≥k0+1. (6.39) Indeed, any interval [ s,t]⊆(0,sk0] can be decomposed as the union of at most 5 intervals whose endpoints fit in one of the 3 possibilities above. Let us prove (6.37). From (6.14) it turns out that |c(ti)−c(tj)|=δ2/vextendsingle/vextendsingle/vextendsingle/vextendsingle/vextendsingle1 λ2−4σ i−1 λ2−4σ j/vextendsingle/vextendsingle/vextendsingle/vextendsingle/vextendsingle≤δ2/vextendsingle/vextendsingle/vextendsingle/vextendsingle1 λ2 i−1 λ2 j/vextendsingle/vextendsingle/vextendsingle/vextendsingle1−2σ , where the inequality follows from the fact that the function x→x1−2σis (1−2σ)-H¨ older continuous with constant equal to 1. Now from (6.12) it follows that δ2/vextendsingle/vextendsingle/vextendsingle/vextendsingle1 λ2 i−1 λ2 j/vextendsingle/vextendsingle/vextendsingle/vextendsingle1−2σ =δ2 (4π)2−4σ|t2 i−t2 j|1−2σ=δ2 (4π)2−4σ|ti+tj|1−2σ|ti−tj|1−2σ ω(|ti−tj|)ω(|ti−tj|). Since|ti+tj| ≤2tk0and|ti−tj| ≤tk0, we conclude that |c(ti)−c(tj)| ≤δ2 (4π)2−4σ(2tk0)1−2σsup/braceleftbiggx1−2σ ω(x):x∈(0,tk0)/bracerightbigg ω(|ti−tj|), so that (6.37) follows from (6.20). Let us prove (6.38). Since c(t) isπ/λkperiodic in [ tk,sk], for every ( a,b)∈[tk,sk]2 there exists ( a1,b1)∈[tk,sk]2such thatc(a) =c(a1),c(b) =c(b1), and|a1−b1| ≤π/λk. Thus from (6.35) it follows that |c(a)−c(b)|=|c(a1)−c(b1)| ≤32εkλk|a1−b1|= 32εkλk|a1−b1| ω(|a1−b1|)ω(|a1−b1|), so that we are left to prove that 32εkλk|a1−b1| ω(|a1−b1|)≤1 5. (6.40) Due to (2.6), (2.5), and the fact that |a1−b1| ≤π/λk, it turns out that |a1−b1| ω(|a1−b1|)≤π/λk ω(π/λk)≤π λkω(1/λk), so that now (6.40) follows from (6.24). Let us prove (6.39). Since c(t) is affine in [ sk,tk−1], for every aandbin this interval it turns out that |c(a)−c(b)|=δ2 tk−1−sk/parenleftbigg1 λ2−4σ k−1−1 λ2−4σ k/parenrightbigg |a−b|. 31Sincesk≤tk−1/2, it follows that |c(a)−c(b)| ≤2δ2 tk−11 λ2−4σ k−1·|a−b|=2δ2 tk−11 λ2−4σ k−1·|a−b| ω(|a−b|)·ω(|a−b|). Due to (2.6), (2.5), and the fact that |a−b| ≤tk−1, it turns out that |a−b| ω(|a−b|)≤tk−1 ω(tk−1)≤tk−1 ω(1/λk−1), so that now (6.39) is a simple consequence of (6.22). Energy functions Let us introduce the classic energy functions Ek(t) :=|u′ k(t)|2+λ2 k|uk(t)|2, Fk(t) :=|u′ k(t)|2+λ2 kc(t)|uk(t)|2. Due to (3.15), they are equivalent in the sense that 1 2Ek(t)≤Fk(t)≤3 2Ek(t)∀t∈R. Therefore, proving (3.17) is equivalent to showing that ∞/summationdisplay k=k0a2 kEk(0)exp(2rϕ(λk))<+∞ ∀r>0, (6.41) while proving (3.18) is equivalent to showing that ∞/summationdisplay k=k0a2 kFk(t)exp(−2Rψ(λk)) = +∞ ∀R>0,∀t>0. (6.42) We are thus left to estimating Ek(0) andFk(t). Estimates in [0,tk] We prove that Ek(0)≤λ2 kexp(4π)∀k≥k0. (6.43) To this end, we begin by estimating Ek(tk). From (6.31) we obtain that uk(tk) = 0 and |u′ k(tk)| ≤λkexp(2εkλktk) =λkexp(8πεk), so that Ek(tk)≤λ2 kexp(16πεk). (6.44) 32Now the time-derivative of Ek(t) is E′ k(t) =−4δλ2σ k|u′ k(t)|2−2λ2 k(c(t)−1)u′ k(t)uk(t)∀t∈R. Therefore, from (3.15) it follows that E′ k(t)≥ −4δλ2σ kEk(t)−λk|c(t)−1|·2|u′ k(t)|·λk|uk(t)| ≥ −/parenleftbigg 4δλ2σ k+λk 2/parenrightbigg Ek(t) for everyt∈R. Integrating this differential inequality in [0 ,tk] we deduce that Ek(0)≤Ek(tk)exp/bracketleftbigg/parenleftbigg 4δλ2σ k+λk 2/parenrightbigg tk/bracketrightbigg . Keeping (6.44) and (6.12) into account, we conclude that Ek(0)≤λ2 kexp/parenleftbigg 16πεk+16πδ λ1−2σ k+2π/parenrightbigg , so that (6.43) follows immediately from (6.18). Estimates in [tk,sk] In this interval it turns out that uk(t) :=w(εk,λk,t), where w(ε,λ,t) is the function defined in (6.1). Keeping (6.14) and (6.15) into accou nt, we obtain that uk(sk) = 0 and |u′ k(sk)|=λkexp(b(εk,λk,sk)) =λkexp/parenleftbig (2εkλk−δλ2σ k)sk/parenrightbig . Therefore, from (6.23) it follows that |u′ k(sk)| ≥λkexp(εkλksk), and hence Fk(sk) =Ek(sk)≥λ2 kexp(2εkλksk). (6.45) Estimates in [sk,tk−1] We prove that Fk(tk−1)≥λ2 kexp(2εkλksk−4δλ2σ ktk−1). (6.46) Indeed the time-derivative of Fk(t) is F′ k(t) =−4δλ2σ k|u′ k(t)|2+λ2 kc′(t)|uk(t)|2∀t∈(sk,tk−1). Sincec′(t)>0 in (sk,tk−1), it follows that F′ k(t)≥ −4δλ2σ k|u′ k(t)|2≥ −4δλ2σ kFk(t)∀t∈(sk,tk−1), and hence Fk(tk−1)≥Fk(sk)exp/parenleftbig −4δλ2σ k(tk−1−sk)/parenrightbig ≥Fk(sk)exp/parenleftbig −4δλ2σ ktk−1/parenrightbig . Keeping (6.45) into account, we have proved (6.46). 33Estimates in [tk−1,+∞) We prove that Fk(t)≥λ2 kexp/parenleftbig 2εkλksk−8δλ2σ kt−64εk−1λk−1t/parenrightbig ∀t≥tk−1.(6.47) To this end, let us set Ik:= [tk−1,+∞)\k−1/uniondisplay i=k0{ti,si}. First of all, we observe that |c′(t)| ≤32εk−1λk−1∀t∈Ik (6.48) Indeed we know from (6.35) and (6.36) that |c′(t)| ≤32εiλi∀t∈(ti,si)∪(si,ti−1), and of course c′(t) = 0 for every t>sk0. Now it is enough to observe that Ik= (tk0,sk0)∪(sk0,+∞)∪k−1/uniondisplay i=k0+1[(ti,si)∪(si,ti−1)], and thatεiλiis a nondecreasing sequence because of (6.30). Now we observe that the function t→Fk(t) is continuous in [ tk−1,+∞) and differ- entiable in Ik, with F′ k(t) =−4δλ2σ k|u′ k(t)|2+λ2 kc′(t)|uk(t)|2 ≥ −4δλ2σ k|u′ k(t)|2−|c′(t)| c(t)·λ2 kc(t)|uk(t)|2 ≥ −/parenleftbigg 4δλ2σ k+|c′(t)| c(t)/parenrightbigg Fk(t). Therefore, from (6.48) and (3.15) it follows that F′ k(t)≥ −/parenleftbig 4δλ2σ k+64εk−1λk−1/parenrightbig Fk(t)∀t∈Ik, and hence Fk(t)≥Fk(tk−1)exp/bracketleftbig −/parenleftbig 4δλ2σ k+64εk−1λk−1/parenrightbig (t−tk−1)/bracketrightbig ≥Fk(tk−1)exp/bracketleftbig −/parenleftbig 4δλ2σ k+64εk−1λk−1/parenrightbig t/bracketrightbig for everyt≥tk−1. Keeping (6.46) into account, we finally obtain that Fk(t)≥λ2 kexp/parenleftbig 2εkλksk−4δλ2σ ktk−1−4δλ2σ kt−64εk−1λk−1t/parenrightbig , from which (6.47) follows by simply remarking that t≥tk−1. 34Conclusion We are now ready to verify (6.41) and (6.42). Indeed from (6.32) an d (6.43) it turns out that a2 kEk(0)exp(2rϕ(λk))≤1 k2λ2 kexp(−2kϕ(λk))·λ2 kexp(4π)·exp(2rϕ(λk)) =1 k2exp(4π+2(r−k)ϕ(λk)). The argument of the exponential is less than 4 πwhenkis large enough, and hence the series in (6.41) converges. Let us consider now (6.42). For every t>0 it turns out that t≥tk−1whenkis large enough. For every such kwe can apply (6.47) and obtain that a2 kFk(t)exp(−2Rψ(λk)) ≥1 k2exp/parenleftbig −2kϕ(λk)−2Rψ(λk)+2εkλksk−8δλ2σ kt−64εk−1λk−1t/parenrightbig . Keeping (6.29) into account, it follows that a2 kFk(t)exp(−2Rψ(λk)) ≥1 k2exp/parenleftbig (k−64t)εk−1λk−1+2(k−R)ψ(λk)+(2k−8δt)λ2σ k+k/parenrightbig ≥1 k2exp(k) whenkis large enough. This proves that the series in (6.42) diverges. /square References [1]G. Chen, D. L. Russell ; A mathematical model for linear elastic systems with structural damping. Quart. Appl. Math. 39(1981/82), no. 4, 433–454. [2]S. P. Chen, R. Triggiani ; Proof of extensions of two conjectures on structural damping for elastic systems. Pacific J. Math. 136(1989), no. 1, 15–55. [3]S. P. Chen, R. Triggiani ; Characterization of domains of fractional powers of certain operators arising in elastic systems, and applications. J. Differential Equa- tions88(1990), no. 2, 279–293. [4]S. P. Chen, R. Triggiani ; Gevrey class semigroups arising from elastic systems with gentle dissipation: the case 0 <α<1/2.Proc. Amer. Math. Soc. 110(1990), no. 2, 401–415. [5]F. Colombini ; Quasianalytic and nonquasianalytic solutions for a class of weakly hyperbolic Cauchy problems. J. Differential Equations 241(2007), no. 2, 293–304. 35[6]F. Colombini, E. De Giorgi, S. Spagnolo ; Sur le ´ equations hyperboliques avec des coefficients qui ne d´ ependent que du temp. (French) Ann. Scuola Norm. Sup. Pisa Cl. Sci. (4) 6(1979), no. 3, 511–559. [7]F. Colombini, E. Jannelli, S. Spagnolo ; Well-posedness intheGevrey classes of the Cauchy problem for a nonstrictly hyperbolic equation with coe fficients de- pending on time. Ann. Scuola Norm. Sup. Pisa Cl. Sci. (4)10(1983), no. 2, 291– 312. [8]M. Ghisi, M. Gobbino ; DerivativelossforKirchhoffequationswithnon-Lipschitz nonlinear term. Ann. Scuola Norm. Sup. Pisa Cl. Sci. (5) 8(2009), no. 4, 613–646. [9]M. Ghisi, M. Gobbino, H. Haraux ; Local and global smoothing effects for some linear hyperbolic equations with a strong dissipation. Trans. Amer. Math. Soc.To appear. Preprint arXiv:1402.6595 [math.AP] . [10]A. Haraux, M. ˆOtani; Analyticity and regularity for a class of second order evolution equation. Evol. Equat. Contr. Theor. 2(2013), no. 1, 101–117. [11]R. Ikehata ; Decayestimatesofsolutionsforthewaveequationswithstrongd amp- ing terms in unbounded domains. Math. Methods Appl. Sci. 24(2001), no. 9, 659– 670. [12]R. Ikehata, M. Natsume ; Energy decay estimates for wave equations with a fractional damping. Differential Integral Equations 25(2012), no. 9-10, 939–956. [13]R. Ikehata, G. Todorova, B. Yordanov ; Wave equations with strong damp- ing in Hilbert spaces. J. Differential Equations 254(2013), no. 8, 3352–3368. [14]J.-L. Lions, E. Magenes , Probl` emes aux limites non homog` enes et applications. Vol. 3. (French) Travaux et Recherches Mathmatiques, No. 20. D unod, Paris, 1970. [15]S. Matthes, M. Reissig ; Qualitative properties of structural damped wave mod- els.Eurasian Math. J. 4(2013), no. 3, 84–106. [16]K. Nishihara ; Degenerate quasilinear hyperbolic equation with strong damping. Funkcial. Ekvac. 27(1984), no. 1, 125–145. [17]K. Nishihara ; Decay properties of solutions of some quasilinear hyperbolic equa- tions with strong damping. Nonlinear Anal. 21(1993), no. 1, 17–21. [18]M. Reed, B. Simon ;Methods of Modern Mathematical Physics, I: Functional Analysis. Second edition . Academic Press, New York, 1980. [19]Y. Shibata ; Onthe rate of decay of solutions to linear viscoelastic equation. Math. Methods Appl. Sci. 23(2000), no. 3, 203–226. 36 | 2014-08-15 | We consider a second order linear equation with a time-dependent coefficient
c(t) in front of the "elastic" operator. For these equations it is well-known
that a higher space-regularity of initial data compensates a lower
time-regularity of c(t).
In this paper we investigate the influence of a strong dissipation, namely a
friction term which depends on a power of the elastic operator.
What we discover is a threshold effect. When the exponent of the elastic
operator in the friction term is greater than 1/2, the damping prevails and the
equation behaves as if the coefficient c(t) were constant. When the exponent is
less than 1/2, the time-regularity of c(t) comes into play. If c(t) is regular
enough, once again the damping prevails. On the contrary, when c(t) is not
regular enough the damping might be ineffective, and there are examples in
which the dissipative equation behaves as the non-dissipative one. As expected,
the stronger is the damping, the lower is the time-regularity threshold.
We also provide counterexamples showing the optimality of our results. | Linear hyperbolic equations with time-dependent propagation speed and strong damping | 1408.3499v1 |
arXiv:1311.3518v1 [math.GT] 14 Nov 2013THE DIMENSION OF THE LEAFWISE REDUCED COHOMOLOGY JES´US A.´ALVAREZ L ´OPEZ AND GILBERT HECTOR Abstract. Geometric conditions are given so that the leafwise reduced co- homology is of infinite dimension, specially for foliations with dense leaves on closed manifolds. The main new definition involved is the int ersection number of subfoliations with “appropriate coefficients”. The leafw ise reduced cohomol- ogy is also described for homogeneous foliations with dense leaves on closed nilmanifolds. 1.Introduction LetFbe aC∞foliation on a manifold M. Theleafwise de Rham complex (Ω·(F),dF) is the restriction to the leaves of the de Rham complex of M; i.e., Ω(F) is the space of differential forms on the leaves that are C∞on the ambient manifoldM, anddFis the de Rham derivative on the leaves. We use the notation Ω(F) =C∞(/logicalandtextT∗F) meaning C∞sections on M. The cohomology H·(F) = H·(Ω(F),dF) is called the leafwise cohomology ofF. It is well known that H·(F) canalso be defined as the cohomologyof Mwith coefficients in the sheafofgerms of C∞functions which are locally constant on the leaves, but we do not use this. The (weak)C∞topology on Ω( F) induces a topology on H·(F), which is non-Hausdorff in general [15]. The quotient space of H·(F) over the closure of its trivial subspace is called the leafwise reduced cohomology ofF, and denoted by H·(F). Similarly, we can also define Ω· c(F),H· c(F) andH· c(F) by considering compactly supported C∞sections of/logicalandtextTF∗. For degree zero we have that H0(F) =H0(F) is the space of C∞functions on Mthat are constant on each leaf—the so called (smooth) basic functions ; thus H0(F)∼=Rif the leaves are dense. Though density of the leaves seems to yield strong restrictions on the leafwise cohomology also for higher degr ee, this cohomol- ogy may be of infinite dimension when leaves are dense and Mis closed. In fact, for dense linear flows on the two-dimensional torus, we have dim H1(F) = 1 when the slope of the leaves is a diophantine irrational number [18], but dim H1(F) =∞ if the slope is a Liouville’s irrational number [30]. Nevertheless H1(F)∼=Rin both cases. This computation was later generalized to the case of linear f oliations on tori of arbitrary dimension [20, 8]. Other known properties of the leafwise cohomology are the following ones. The leafwisecohomologyofdegreeonewithcoefficientsinthenormalbun dleisrelatedto the infinitesimal deformations of the foliation [18]. For p= dimF, the dual space Hp c(F)′is canonically isomorphic to the space of holonomy invariant transver se distributions [15]—recall that for a topological vector space V, the dual space V′is 1991Mathematics Subject Classification. 57R30. Partially supported by Xunta de Galicia (Spain). 12 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR the space of continuous linear maps V→R.H·(F) is invariant by leaf preserving homotopies, and Mayer-Vietoris arguments can be applied [10], whic h was used to computeH·(F) for someexamples. For an arbitraryflow Fon the two-torus, it was proved that dim H1(F) =∞ifFis not minimal, and dim H1(F) = 1 if and only if FisC∞conjugate to a Diophantine linear flow [7]. The triviality of H1(F) implies the triviality of the linear holonomy [10], and is equivalent to Thurston’s stability if codim F= 1 andMis closed [6]. However, more general relations between the leafwise cohomology and the geometry of the foliation remain rather unknown. The above examples of linear foliations on tori could wrongly suggest thatH·(F) maybe offinite dimension if Mis closedand the leavesaredense. In fact, S. Hurder and the first author gave examples of foliations with dense leaves on closed Rie- mannian manifolds with an infinite dimensional space of leafwise harmon ic forms that areC∞on the ambient manifold [4], and this space is canonically injected in the leafwise reduced cohomology; indeed this injection is an isomorph ism at least for the so called Riemannian foliations [5]. So a natural problem is the fo llowing: Give geometric properties characterizing C∞foliations whose leafwise reduced co- homology is of finite dimension; specially for foliations wi th dense leaves on closed manifolds . The aim of this paper is to give an approach to this problem. The first a nd main geometric idea we use is the intersection number of subfoliations with “appropriate coefficients”. To explain it, consider the simplest example where M=T×Lwith thefoliation Fwhoseleavesarethe slices {∗}×L,whereT,Lareclosedmanifolds of dimensions q,p. Let (Ω·(L),dL) be the de Rhamcomplex of L, and letH·(L),H·(L) denote the homology and cohomology of Lwith real coefficients. Then Ω·(F) is theC∞closure ofC∞(T)⊗Ω·(L), wheredF= 1⊗dL. So Hk(F) =Hk(F)≡C∞(T)⊗Hk(L) becauseH·(L) is of finite dimension. Assume Lis oriented for simplicity. Then recall that Poincar´ e duality and integration of differential forms e stablish canonical isomorphisms Hk(L)∼=Hq−k(L), Hk(L)′∼=Hk(L), such that the canonical pairing between Hk(L)⊗Hk(L)′→Rcorresponds to the intersection pairing Hp−k(L)⊗Hk(L)→R[32]. Hence Hk(F)∼=C∞(T)⊗Hp−k(L), (1) Hk(F)′∼=C∞(T)′⊗Hk(L), (2) such that the canonical pairing Hk(F)⊗Hk(F)′→Rcorresponds to the product of the evaluation of distributions on C∞functions and the intersection pairing. Now observe that, according to [32], the right hand side spaces in (1 ) and (2) are respectively generated by elements of the form f⊗[K1] andD⊗[K2], wherefis aC∞function on T,Dis a distribution on T, andK1,K2⊂Lare closed oriented submanifolds of dimensions p−k,k. Hence dim Hk(F) =∞is equivalent to the existence of sequences of elements fm⊗[K1] andDn⊗[K2] as above so that K1,K2 havenon-trivial intersection number and Dn(fn)/\e}atio\slash= 0 if and only if m=n; of course this holds just when Hk(L)/\e}atio\slash= 0 andq>0. Now consider each element fm⊗[K1] as the family of homology classes fm(t)[{t}×K1]∈Hp−k({t}×L), t∈T ,LEAFWISE REDUCED COHOMOLOGY 3 determined by the family of closed oriented submanifolds {t} ×K1of the leaves ofFand the family of coefficients fm(t). The elements Dn⊗[K2] have a similar interpretationby considering distributions as generalized function s. A key property here is that the families {t}×K1and{t}×K2depend smoothly on t, determining C∞subfoliations F1,F2ofF. Other key properties are the C∞dependence of the coefficients fm(t) ont, and the distributional dependence of the generalized coefficients Dn(t) ont. This means that the fmareC∞basic functions of F1 and theDnare “distributional basic functions” of F2; i.e., theDnare holonomy invariant transverse distributions of F2. It turns out that these key properties are enough to generalize the above ideas in a way that can be applied even when the leavesaredense, obtainingourfirstmaintheoremthatroughlyass ertsthefollowing: For aC∞oriented foliation Fof dimension p, we have dimHk c(F) =∞whenF has oriented subfoliations F1,F2of dimensions k−p,p, and there is a sequence of basic functions fmofF1and a sequence of transverse invariant distributions Dnof F2, such that the corresponding “intersection numbers” are no n-trivial if and only if m=n—certain simple conditions are also required for the “inter section numbers” to be defined . We do not know whether such conditions form a characterization of the cases where dim Hk c(F) =∞; this depends on whether it is possible to “smoothen” the representatives of classes in certain leafwise hom ologies introduced in [3]. Indeed the above fmandDnplay the rˆ ole of coefficients in homology, assigning a number to each leaf of the subfoliations; the way these n umbers vary from leaf to leaf is what makes these coefficients appropriate. Though these conditions are difficult to check in general, this result h as many easy to apply corollaries. For instance, suppose an oriented foliatio nFisRiemann- ian—in the sense that all of its holonomy transformations are local isom etries for some Riemannian metric on local transversals [29, 25]. Then dim H· c(F) =∞if Fis of positive codimension and some leaf of Fcontains homology classes with non-trivial intersection. These conditions are quite simple to verify . In this case, the infinitely many linearly independent classes obtained in H· c(F) can be consid- ered as “transverse diffusions” of the homology classes in the leaf. This diffusion idea is inspired by the unpublished preprints [19, 4]. Indeed [19] is the g erminal work about the relation of the analysis on the leaves and on the ambie nt manifold obtained by transverse diffusion. Other consequences of the above general theorem hold when Fis asuspension foliation. That is, the ambient manifold of Fis the total space of a fiber bundle M→Bwith the leaves transverse to the fibers, and such that the restr iction of the bundle projection to each leaf is a regular covering of the bas eB. Now dimH· c(F) =∞whenBis oriented and has homology classes with non-trivial intersection satisfying additional properties with respect to the h olonomy of F. In this case the leaves may not contain homology classes with non-trivia l intersection, and thus the idea of “transverse diffusion” of homology classes in th e leaves may fail. In fact we shall see that the infinite dimension of H1 c(F) may be more related to the number of ends of the leaves. To explain another theorem of this paper, recall that a foliation Fon a manifold Mis aLie foliation when it has a complete transversal diffeomorphic to an open subset of a Lie group Gso that holonomy transformationson this transversalcorre- spond to restrictions of left translations on G—this type of foliations play a central rˆ ole in the study of Riemannian foliations [25]. The Lie algebra gofGis called the4 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR structural Lie algebra ofF; we may also simply say that Fis a Lieg-foliation. In this case, if Mis closed and oriented, and gis compact semisimple, then we obtain that dim H·(F) =∞when some additional hypotheses are satisfied. Again we use homology classes with non-trivial intersection in the hypotheses, b ut now they live in the homology of M. The proof of this result is reduced to the case of suspension foliations to apply what we already know. This reduction process con tains rather delicate arguments based on the work [2] of the first author. The above results are negative in the sense that all of them give con ditions for the nonexistence of finite Betti numbers for the reduced leafwise cohomology. In contrast, our final theorem shows that the reduced leafwise coh omology of the so calledhomogeneous foliations with dense leaves in closed nilmanifolds is isomorphic to the cohomology of the Lie algebra defining the foliation. This has be en also proved by X. Masa with different techniques. Acknowledgment . We wish to thank F. Alcalde for many helpful conversations. The first author would like to thank the hospitality of the Institut de Math´ ema- tiques et d’Informatique of the University Claude Bernard of Lyon s everal times during the preparation of this work. We would like also to thank the re feree for important corrections. 2.Main results For the sake of simplicity, all manifolds, foliations, maps, functions, differential forms and actions will be assumed to be C∞from now on, unless the contrary is explicitly stated. LetFbe a foliation on a manifold M. For any closed saturated subset S⊂M, let Ω· S(F)⊂Ω·(F) be the subcomplex of leafwise differential forms whose support has compact intersection with S. Consider the topology on Ω· S(F) determined as follows: A sequence αn∈Ω· S(F) converges to zero if it converges to zero in Ω·(F) and there is a compact subset K⊂Ssuch thatS∩suppαn⊂Kfor alln. We have the corresponding cohomology H· S(F), and reduced cohomology H· S(F). With this notations, observe that Ω·(F) = Ω· ∅(F) and Ω· c(F) = Ω· M(F) as topological vector spaces. Letf: (M1,F1)→(M2,F2) be a map of foliated manifolds, and let Si⊂Mi, i= 1,2, beclosedsaturatedsubsetssuchthat the restriction f:S1→S2is aproper map. Then f∗(Ω· S2(F2))⊂Ω· S1(F1), yielding a homomorphism f∗:H· S2(F2)→ H· S1(F1). In particular we get f∗:H· c(F2)→ H· S1(F1) iff:S1→M2is proper. The following is what we need to define the intersection number of sub foliations with “appropriate coefficients”: •An oriented foliation Fon a manifold M, and two immersed oriented sub- foliationsιi: (Mi,Fi)→(M,F),i= 1,2. •dimF= dimF1+dimF2, and codim F= codim F1. •Eachιiistransversely regular in the sense that it defines embeddings of small enough local transversals of Fiinto local transversals of F; i.e. the homomorphism defined by the differential of ιibetween the normal bundles ofFiandFis injective on the fibers. •A compactly supported basic function fofF1. •A holonomy invariant transverse distribution DofF2such that the map ι2: suppD→Mis proper.LEAFWISE REDUCED COHOMOLOGY 5 •The restrictions ι1|suppfandι2|suppDintersect transversely inFin the sense that, for all leaves LiofFiandLofFsuch thatL1⊂suppf,L2⊂ suppDandι1(L1)∪ι2(L2)⊂L, the immersed submanifolds ιi:Li→L intersect transversely in L. Observe that there are open neighborhoods, N1of suppfandN2of suppD, such that theιi|Niintersect transversely in F. Consider the pull-back diagram Tσ1− −−− →N1 σ2/arrowbt/arrowbtι1 N2ι2− −−− →M . Here T={(x1,x2)∈N1×N2|ι1(x1) =ι2(x2)}, and theσiare restrictions of the factor projections. It is easy to check th atι1×ι2: N1×N2→M×Mis transverse to the diagonal ∆, and thus Tis a manifold with dimT= codim F2. Moreover the σiare immersions, and σ2is transverse toF2. SoDdefines a distribution on T, which will be denoted by DT. We also have the locally constant intersection function ε:T→ {±1}, whereε(x1,x2) =±1 depending on whether the identity Tιi(xi)F ≡ι1∗Tx1F1⊕ι2∗Tx2F2 is orientation preserving or orientation reversing. On the other ha nd (ι1(suppf)×ι2(suppD))∩∆ iscompactbecauseit isaclosedsubsetofthe compactspace ι1(suppf)×ι1(suppf). So suppσ∗ 1f∩suppDT= (ι1×ι2)−1((ι1(suppf)×ι2(suppD))∩∆) is a compact subspace of Tsinceι1×ι2: suppf×suppD→M×Mis a proper map. Thus the following definition makes sense. Definition 2.1. With the above notations, the intersection number of (ι1,f) and (ι2,D), denoted by /a\}b∇acketle{t(ι1,f),(ι2,D)/a\}b∇acket∇i}ht, is defined as DT(g) for any compactly sup- ported function gonTwhich is equal to the product εσ∗ 1fon some neighborhood of suppσ∗ 1f∩suppDT. Now our first main theorem is the following. Theorem 2.2. LetFbe an oriented foliation on a manifold M, andιi: (Mi,Fi)→ (M,F),i= 1,2, transversely regular immersed oriented subfoliations. S uppose dimF= dimF1+ dimF2, andcodimF= codim F1. Letfmbe a sequence of compactly supported basic functions of F1, andDna sequence of holonomy invari- ant transverse distributions of F2such that each restriction ι2: suppDn→Mis a proper map. Suppose each pair ι1|suppfmandι2|suppDnintersect transversely in F, and/a\}b∇acketle{t(ι1,fm),(ι2,Dn)/a\}b∇acket∇i}ht /\e}atio\slash= 0if and only if m=n. ThendimHk c(F) =∞for k= dimF2. The following two corollaries are the first type of consequences of T heorem 2.2; the second corollary follows directly from the first one.6 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR Corollary 2.3. LetFbe an oriented foliation of codimension q >0,La leaf of F, andh:π1(L)→G∞ qits holonomy representation, where G∞ qis the group of germs at the origin of local diffeomorphisms of Rqwith the origin as fixed point. Let ιi:Ki→L,i= 1,2, be smooth immersions of closed oriented manifolds of com- plementary dimension and nontrivial intersection. Suppos e there is a Riemannian metric on Rqso that the elements in the image of the composites (3) π1(Ki)π1(ιi)− −−− →π1(L)h− −−− →G∞ q are germs of local isometries. Then dimHki c(F) =∞forki= dimKi,i= 1,2. Corollary 2.4. LetFbe an oriented Riemannian foliation of positive codimensio n. Suppose some leaf of Fhas homology classes of complementary degrees, k1andk2, with non-trivial intersection. Then dimHki c(F) =∞,i= 1,2. BeforestatingthenexttypeofcorollariesofTheorem2.2, recallt hatasuspension foliation Fis given as follows. Let π:L→Bbe a regular covering map of an oriented manifold, and let Γ be its group of deck transformations. F or any effective action of Γ on some manifold T, consider the right diagonal action of Γ on L×T: (z,t)γ= (zγ,γ−1t) forγ∈Γ and (z,t)∈L×T. Then Fis the foliation on M= (L×T)/Γwhoseleavesaretheprojectionsofthesubmanifolds L×{∗} ⊂L×T. The element in Mdefined by each ( z,t)∈L×Twill be denoted by [ z,t]. The mapρ:M→Bgiven byρ([z,t]) =π(z) is a fiber bundle projection with typical fiberT. The leaves of Fare transverse to the fibers of ρ, and define coverings of B. The leaf that contains [ z,t] can be canonically identified to L/Γt, where Γ tis the isotropy subgroup of Γ at t. This leaf is dense if and only if the Γ-orbit of tis dense inT. Corollary 2.5. With the above notation, let h:π1(B)→Γbe the surjective homomorphism defined by the regular covering LofB, and letιi:Ki→B,i= 1,2, be immersions of connected oriented manifolds of complemen tary dimension in B. SupposeK1is a closed manifold, ι2a proper map, and the homology class defined byι1has non-trivial intersection with the locally finite homolo gy class defined by ι2. For each i, letΓi⊂Γbe the image of the composite π1(Ki)π1(ιi)− −−− →π1(B)h− −−− →Γ. Letfmbe a sequence of compactly supported Γ1-invariant functions on T, andDn a sequence of Γ2-invariant distributions on Tsuch thatDn(fm)/\e}atio\slash= 0if and only if m=n. ThendimHk c(F) =∞fork= dimK2. Corollary 2.6. LetB,L,h,Γ,T,F,Ki,ιiandΓibe as in Corollary 2.5. Let µbe aΓ2-invariant measure on T. Suppose the closure of the image of Γ1in the topological group of diffeomorphisms of T(with the weak C∞topology)is a compact Lie group, and there is an infinite sequence of Γ1-invariant open subsets of Twith non-trivial µ-measure and pairwise disjoint Γ2-saturations. Then dimHk c(F) =∞ fork= dimK2. Observe that, in Corollary 2.6, the infinite sequence of Γ 1-invariant open sets may not be Γ 2-invariant, and their Γ 2-saturations may not be Γ 1-invariant. Corollary 2.7. LetB,L,h,Γ,TandFbe as in Corollary 2.5. Suppose that there is a loopc:S1→Bwith a lift to Lthat joins two distinct points of the end set of L. Leta=h([c])∈Γ, where[c]is the element of π1(B)represented by c, and assumeLEAFWISE REDUCED COHOMOLOGY 7 that the closure Hof the image of /a\}b∇acketle{ta/a\}b∇acket∇i}htin the topological group of diffeomorphisms ofT(with the weak C∞topology)is a compact Lie group. Suppose also that there is an infinite sequence of disjoint non-trivial H-invariant open subsets of T. Then dimH1 c(F) =∞. In Corollaries 2.5, 2.6 and 2.7, if Bis compact, then the leaves of Fcan only be dense when Lhas either one end or a Cantor space of ends, as follows from the following. Proposition 2.8. LetΓbe a finitely generated group with two ends, and C⊂Γ an infinite subgroup. Suppose Γacts continuously on some connected T1topological spaceX. Then the Γ-orbits are dense in Xif and only if so are the C-orbits. Now letFbe a Lie g-foliation on a closed manifold M. The following property characterizes such a type of foliations [11, 24, 25]. Let /tildewiderMbe the universal covering ofM,/tildewideFthe lift of Fto/tildewiderM, andGthe simply connected Lie group with Lie algebrag. Then the leaves of /tildewideFare the fibers of a fiber bundle /tildewiderM→G, which is equivariant with respect to some homomorphism h:π1(M)→G, where we consider the right action of π1(M) on/tildewiderMby deck transformations and the right action ofGon itself by right translations. This hand its image are respectively called the holonomy homomorphism andholonomy group ofF. Observe that the fibers ofDare connected because Gis simply connected (a connected coveringof G is given by the quotient of /tildewiderMwhose points are the connected components of these fibers). Theorem 2.9. With the above notation, suppose that Mis oriented and the struc- tural Lie algebra gofFis compact semisimple. Let ιi:Ki→M,i= 1,2, be immersions of closed oriented manifolds of complementar y dimension defining homology classes of Mwith non-trivial intersection. Let Γibe the image of the composite π1(Ki)π1(ιi)− −−− →π1(M)h− −−− →G. Suppose the group generated by Γ1∪Γ2is not dense in G. Letk= dimK2, and suppose either 1≤k≤2orι1is transverse to F. ThendimHk(F) =∞. The following is our final theorem. Theorem 2.10. LetHbe a simply connected nilpotent Lie group, K⊂Ha normal connected subgroup, and Γ⊂Ha discrete uniform subgroup whose projection to H/Kis dense. Let Fbe the foliation of the closed nilmanifold Γ\Hdefined as the quotient of the foliation on Hwhose leaves are the translates of K. Then there is a canonical isomorphism H·(F)∼=H·(k), wherekis the Lie algebra of K. The following two examples are of different nature. In both of them t here are infinitely many linearly independent leafwise reduced cohomology class es of degree one. But these classes are induced by the handles in the leaves in Exa mple 2.11, whereasthey areinduced by the “branches”of the leavesthat de fine a Cantorspace of ends in Example 2.12. Example 2.11 ([4]).LetLbe aZ-covering of the compact oriented surface of genus two; i.e., Lis a cylinder with infinitely many handles attached to it. Each handle contains two circles defining homology classes with non-trivial intersection. Hence for any injection of Zinto then-torusRn/Zn, the corresponding suspension8 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR foliation fulfills the hypotheses of Corollary 2.6, and thus has infinite d imensional reduced leafwise cohomologyof degree one. We could also use Corolla ry 2.4 instead of Corollary 2.6. Example 2.12. Let Γ be the free group with two generators, and La Γ-covering of the compact orientable surface of genus two. This Lhas a Cantor space of ends. Hence, for any injective homomorphism of Γ in a compact Lie group, t he reduced leafwise cohomology of degree one of the corresponding suspensio n foliation is of infinite dimension by Corollary 2.7. 3.Leafwise reduced cohomology and subfoliations This section is devoted to the proof of Theorem 2.2. With the notatio ns in- troduced in Section 2, the idea of the proof is the following. The ( ι1,fm) yield elementsζm∈ Hr c(F) by “leafwise Poincar´ eduality”. On the other hand, the argu- mentsin[15]showthateach Dncanbeconsideredasanelementin Hr Sn(F2)′, where Sn= suppDn⊂M. Moreover there are homomorphisms ι∗ 2:H· c(F)→ H· Sn(F2) since theι2:Sn→Mare proper maps. Then the result follows by verifying /a\}b∇acketle{t(ι1,fm),(ι2,Dn)/a\}b∇acket∇i}ht=Dn(ι∗ 2ζm). We first explain the way “leafwise Poincar´ e duality” works. Consider thetrans- verse complex Ω· c(TrF) introduced in [15], which will be only used for degree zero. For any representative Hof the holonomy pseudogroup of F, defined on some manifoldT, Ω0 c(TrF) is defined as the quotient of C∞c(T) over the subspace gen- erated by the functions of the type φ−h∗φ, whereh∈ Handφ∈C∞c(T) with suppφ⊂domh. As a topological vector space, Ω0 c(TrF) is independent of chosen representative of the holonomy pseudogroup. From the definition it easily follows that the dualspaceΩ0 c(TrF)′canbe canonicallyidentified tothe spaceofholonomy invariant transverse distributions of F. Now consider the representative Hof the holonomy pseudogroup induced by an appropriate locally finite covering of Mby foliation patches Ui; that is, iffi:Ui→ Tiis the localquotient map whosefibers arethe plaques in Ui, then appropriateness of this covering means that each equality fj=hi,jfionUi∩Ujdetermines a diffeomorphisms hi,j:fi(Ui∩Uj)→fj(Ui∩Uj), and the collection of all of these diffeomorphisms generate the pseudogroup HonT=/unionsqtext iTi. Fix also a partition of unityφisubordinated to the covering Ui. With these data we have a map Ωp c(F)→Ω· c(T) given by α/mapsto→/summationtext i/integraltext fiφiα, wherep= dimFand/integraltext fidenotes integration along the fibers of fi. This “integration along the leaves” induces an isomorphism Hp(F)∼=Ω0 c(TrF) of topological vector spaces, which is independent of the choice of the Uiandφi[15,§3.3]. So Hp c(F)′≡Hp c(F)′∼=Ω0 c(TrF)′; i.e., any holonomy invariant distribution Dcan be canonically considered as an element in Hp c(F)′. Moreover Dcan be also considered as an element in Hp S(F)′≡ Hp S(F)′forS= suppDasfollowsfromthe followingargument. Forany α∈Ωp c(F), it is easily verified that D/parenleftBig/summationtext i/integraltext fiφiα/parenrightBig depends only on the restriction of αto any neighborhood of the support of DinM. Therefore, if ζ∈Hp S(F),α∈Ωp S(F) is any representative of ζ, andβ∈Ωp c(F) has the same restriction as αto some neighborhood of S, thenD/parenleftBig/summationtext i/integraltext fiφiβ/parenrightBig does not depend on the choices of αand β, and thus this is a good definition of D(ζ).LEAFWISE REDUCED COHOMOLOGY 9 Theorem 2.2 will follow easily from the following result, which will be prove d in Section 4. Proposition 3.1. LetFbe an oriented foliation on a manifold M. Letι1: (M1,F1)→(M,F)be a transversely regular immersed oriented subfoliation w ith codimF= codim F1, andfa compactly supported basic function of F1. Then there is a class ζ∈ Hk c(F),k= dimF −dimF1, such that (4) /a\}b∇acketle{t(ι1,f),(ι2,D)/a\}b∇acket∇i}ht=D(ι∗ 2ζ) for any subfoliation ι2: (M2,F2)→(M,F)and any holonomy invariant transverse distribution DofF2so that the left hand side of (4)is defined. In the right hand side of(4),Dis considered as an element of Hk S(F2)′forS= suppD, andι∗ 2 denotes the homomorphism Hk c(F)→ Hk S(F2)induced byι2, which is defined since ι2:S→Mis a proper map. We do not know whether (4) completely determines ζ. If so,ζcould be called theleafwise Poincar´ e dual class of (ι1,f). Proof of Theorem 2.2. Letζm∈ Hk c(F) be the classes defined by the ( ι1,fm) ac- cording to Proposition 3.1. If Pn∈ Hk c(F)′is given by the composite Hk c(F)ι∗ 2− −−− → Hk Sn(F2)Dn− −−− →R, we havePn(ζm)/\e}atio\slash= 0 if and only if m=nby Proposition 3.1, yielding the linear independence of the ζm. /square 4.Leafwise Poincar ´e duality This section will be devoted to the proof of Proposition 3.1. 4.1.On the Thom class of a vector bundle. Thefollowinglemma isatechnical step in the proof of Proposition 3.1, which will be proved in Section 4.2. Lemma 4.1. LetMbe a manifold and π:E→Man oriented vector bundle. IdentifyMto the image of the zero section, whose normal bundle is canon ically oriented. There is a sequence Φnof representatives of the Thom class of Esuch that, iffis any function on M, V is any neighborhood of MinE,K⊂Mis any compact subset, and φ:V→Eis any map which restricts to the identity on M and its differential induces an orientation preserving auto morphism of the normal bundle ofM, thenπ−1(K)∩φ−1(suppΦ n)is compact for large enough n, and the sequence of functions/integraltext πφ∗(π∗fΦn)converges to foverKwith respect to the C∞ topology. Corollary 4.2. Letπ:E→Mbe an oriented vector bundle, and ι:N→Man immersion. Let πN:ι∗E→Nbe the pull-back vector bundle, and ˜ι:ι∗E→E the canonical homomorphism. Identify MandNto the image of the zero sections ofEandι∗E, respectively, and consider the induced orientations on th eir normal bundles. Let Vbe an open neighborhood of Ninι∗E, andh:V→Ean extension ofιsuch that the homomorphism between the normal bundles of NandM, defined by the differential of h, restricts to orientation preserving isomorphisms betwee n the fibers. Let Φnbe the forms on Egiven by Lemma 4.1, K⊂Na compact subset, and fa function on M. Thenπ−1 N(K)∩h−1(suppΦ n)is compact for large enoughn, and the sequence of functions/integraltext πNh∗(π∗fΦn)converge to ι∗foverK with respect to the C∞topology.10 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR Proof.LetU1,...,U mbe a finite open cover of Ksuch that each ι:Ui→Mis an embedding. For each i, there is a compactly supported function fionMwhich is supportedinsometubularneighborhood Wiofι(Ui), andsuchthat f=f1+···+fm on some neighborhood of ι(K). Then, taking a neighborhood Viof eachUiinVso thath:Vi→Eis an embedding, we get (5)/integraldisplay πNh∗(π∗fΦn) =/summationdisplay i/integraldisplay πN|π−1 N(Ui)(h|Vi)∗(π∗fiΦn) aroundK, yielding the result if each term in the right hand side of (5) converge s toι∗fi. Therefore we can assume ι, ˜ιandhare embeddings. With this assumption, there is an open disk bundle DoverV, and extensions ˜ι′,h′:D→π−1(V) of ˜ιandh, respectively, which are diffeomorphisms onto open subsets ofE. Letφdenote the composite ˜ι′(D)(˜ι′)−1 − −−− →Dh′ − −−− →π−1(V). Clearly,φsatisfies the conditions of Lemma 4.1, and we can suppose fis supported in ˜ι′(D). So/integraltext πφ∗(π∗fΦn) converges to fover any compact subset of Vwith respect to the C∞topology. But ι∗/integraldisplay πφ∗(π∗fΦn) =/integraldisplay πN((˜ι′)∗φ∗(π∗fΦn)|V) =/integraldisplay πNh∗(π∗fΦn), and the result follows. /square Observe that Lemma 4.1 is a particular case of Corollary 4.2. The coro llary could be proved directly with the arguments of the lemma, but the no tation would become more complicated. 4.2.Proof of Lemma 4.1. The following easy observations will be used to prove Lemma 4.1. Remark 1.LetEandFbe vector bundles over the manifolds MandN, respec- tively. Suppose f:E→Fis a homomorphism which restricts to isomorphisms on the fibers, and let g:M→Nbe the map induced by f. Thus the homomor- phismE→g∗F, canonically defined by f, is an isomorphism. Therefore there is a composite of homeomorphisms C∞(F)→C∞(g∗F)→C∞(E). Here, the first homomorphism is canonically defined by the pull-back d iagram of g∗F, and the second one is induced by the inverse of E→g∗F. Ifs/mapsto→s′by the above composite, then s′is determined by f(s′(x)) =s(g(x)) forx∈M. Remark 2.SetE=Rn×Rk, and letπi,i= 1,2, denote the factor projections of EontoRnandRk, respectively. Let Kbe a compact subset of Rn, andφ:V→W a diffeomorphism between open neighborhoods of Rn×{0}. Supposeφrestricts to the identity on Rn× {0}. For anyr >0, letBr,Sr⊂Rkrespectively denote the Euclidean ball and the Euclidean sphere of radius r centered at the o rigin. Then there is anR>0 and an open neighborhood UofKsuch that, for every x∈UandLEAFWISE REDUCED COHOMOLOGY 11 everyy∈BR,{x} ×Rkintersects transversely φ−1(Rn× {y}) at just one point. Moreover, the map σ:U×BR→(U×Rk)∩φ−1(Rn×BR), determined by {σ(x,y)}= ({x}×Rk)∩φ−1(Rn×{y}), is a diffeomorphism. Indeed σis smooth because each ( U×Rk)∩φ−1(Rn×{y}) can be given as the graph of a map ψy:U→Rkdepending smoothly on y∈BR, and σ(x,y) = (x,ψy(x)). It also has a smooth inverse since ( x,y) = (x,π2φσ(x,y)). Therefore, for r≤R,π1: (U×Rk)∩φ−1(Rn×Sr)→Uis a sphere bundle, whose fibers are of volume uniformly bounded by Crk−1for someC >0 ifUandRare small enough. To begin with the proof of Lemma 4.1, fix a Riemannian structure on E, and letBr,Sr⊂Erespectively denote the corresponding open disk bundle and spher e bundle of radius r. SetS=S1. Letψbe a global angular form of S[9,§11]. (IfE is of rankk,ψis a differential form of degree k−1 restricting to unitary volume forms on the fibers and so that dψ=−π∗e, whereerepresents the Euler class of S.) Letr:E→Rdenote the radius function, and h:E\M→Sthe deformation retraction given by h(v) =v/r(v). For each n, let alsoρnbe a function on [0 ,∞) such that −1≤ρn≤0,ρ′ n≥0,ρn≡ −1 on a neighborhood of 0, and ρn≡0 on [1/n,∞). Then each Φn=d(ρn(r)h∗ψ) =ρ′ n(r)dr∧h∗ψ−ρn(r)π∗e represents the Thom class of E[9,§12]. Local orthonormal frames canonically define isomorphisms of trivia lity ofE which restrict to local isomorphisms between restrictions of each Srand trivial sphere bundles with typical fiber the Euclidean sphere of radius r. So Remark 2 and the conditions satisfied by φyield the existence of some R,C >0 and some relatively compact open neighborhood UofKinMso that •π−1(U)∩φ−1(BR)⊂V, •the map φ:π−1(U)∩φ−1(BR)→φπ−1(U)∩BR is a diffeomorphism whose differential is of fiberwise uniformly bounded norm, and •for 0< r≤R,φ−1(Sr) is transverse to the fibers of πoverUandπ: π−1(U)∩φ−1(Sr)→Uis a sphere bundle whose fibers are of volume uniformly bounded by Crk−1. Theφ∗Φnalso represent the Thom class of EoverUforn>1/R. Hence f−/integraldisplay πφ∗(π∗fΦn) =/integraldisplay π(π∗f−φ∗π∗f)φ∗Φn =/integraldisplay1/n 0ρ′ n(r)dr/integraldisplay π|π−1(U)∩φ−1(Sr)(π∗f−φ∗π∗f)φ∗h∗ψ (6) −/integraldisplay π(π∗f−φ∗π∗f)ρn(φ∗r)φ∗π∗e. (7)12 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR We have to prove that (6) and (7) converge uniformly to zero on Kasn→ ∞, as well as all of its derivatives of any order. Take a Riemannian metric on M, and a splitting TE=V ⊕H, whereVis the vertical bundle of πandHthe horizontal bundle of any Riemannian connection. This yields a Riemannian structure on TEdefined in the standard way by using the canonical isomorphisms V∼=π∗EandH∼=π∗TM. We also have TM=H|M, H|Sr⊂TSr, and (8) TSr= (V ∩TSr)⊕(H|Sr). Finally, we can assume (9) V ∩φ−1 ∗(H) = 0 over π−1(U)∩φ−1(BR) by the properties of φ. By the conditions on φ, the supremum of |π∗f−φ∗π∗f|overπ−1(U)∩φ−1(B1/n) convergestozeroas n→ ∞. Alsothepointwisenormof φ∗π∗eisuniformlybounded onπ−1(U)∩φ−1(B1/n), thus (7) converges uniformly to zero as n→ ∞. On the otherhand, because the fiberwise normof each h∗:TSr→TSisr−1, the pointwise norm ofφ∗h∗ψis uniformly bounded on π−1(U)∩φ−1(Sr) byC1r−k+1withC1>0 independent of r≤R. So (6) also converges uniformly to zero on Uasn→ ∞by the estimate on the volume of the fibers of πonπ−1(U)∩φ−1(Sr). Now fix vector fields X1,...,X monU. By (8) and (9) the Xihave liftings Yi which are sections of φ−1 ∗Hoverπ−1(U)∩φ−1(BR). For any subset I⊂ {1,...,m}, letθIdenote the composite of Lie derivatives θY1···θYlifI={i1,...il}with i1< i2<···< il, and letθ∅be the identity homomorphism. Then the order m derivativeX1···Xmover (6) and (7) is respectively given by (10)/summationdisplay I,J/integraldisplay1/n 0ρ′ n(r)dr/integraldisplay π|π−1(U)∩φ−1(Sr)θI(π∗f−φ∗π∗f)θJφ∗h∗ψ, and (11) −/summationdisplay I,J/integraldisplay πθI(π∗f−φ∗π∗f)ρn(φ∗r)θJφ∗π∗e, whereI,Jruns over the partitions of {1,...,m}. By the properties of Handφ, the supremum of the |θI(π∗f−φ∗π∗f)|onπ−1(U)∩φ−1(B1/n) converges to zero asn→ ∞. Hence (11) converges uniformly to zero on Kbecause the pointwise normofthe θJφ∗π∗ecanbe uniformlybounded on π−1(K)∩φ−1(BR). The uniform convergenceof (10)tozerofollowsbyestimatingthepointwisenor moftheθJφ∗h∗ψ onπ−1(K)∩φ−1(Sr) byC2r−k+1for someC2>0 independent of r. This in turn follows by proving a similar estimate for the pointwise norm of θ′ Jh∗ψon φπ−1(K)∩Sr, where the θ′ Jare defined in the same way as the θJby using the Y′ i=φ∗Yiinstead of the Yi. To do this, consider the multiplication map µ: [0,R]×S→BR. Since µ∗: [0,1]×(H|S)⊂T([0,1]×S)→ H restricts to isomorphisms on the fibers, by Remark 1 there are smo oth sections Y′′ i of [0,1]×(H|S) so thatµ∗(Y′′ i(r,v)) =Y′ i(rv). Also because the composite (0,R]×Sµ− −−− →BR\Mh− −−− →SLEAFWISE REDUCED COHOMOLOGY 13 is the second factor projection, µ∗h∗ψis the form canonically defined by ψon (0,R]×S, which extends smoothly to [0 ,R]×S. So, ifθ′′ Jis defined in the same way as theθJby using the Y′′ iinstead of the Yi, the pointwise norm of the θ′′ Jµ∗h∗ψ is uniformly bounded. Then the desired estimation of the pointwise no rm of the θ′ Jh∗ψfollows by observing that the fiberwise norm of µ∗:{r}×TS→TSrisr. 4.3.Proof of Proposition 3.1. Recall that any local diffeomorphism φ:M→N induces a homomorphism of complexes, φ∗: Ωc(M)→Ωc(N), defined as follows. For anyα∈Ωc(M), choose a finite open cover U1,...,U nof suppαsuch that each restriction φ:Ui→φ(Ui) is a diffeomorphism. There is a decomposition α=α1+···+αnso that supp αi⊂Ui. For eachi, there is a unique βi∈Ωc(N) supported in φ(Ui) such that βi|φ(Ui)corresponds to αi|Uibyφ. Defineφ∗α= β1+···+βn. This definition is easily checked to be independent of the choices involved and compatible with the differential maps. If φ: (M,F)→(N,G) is a local diffeomorphism of foliated manifolds, we similarly have a homomor phism φ∗: Ωc(F)→Ωc(G) which is compatible with the leafwise de Rham derivative. Moreoverφ∗is surjective if so is φ. Now Proposition 3.1 can be proved as follows. There is a canonical injection of TF1as vector subbundle of ι∗ 1TF. LetE= ι∗ 1TF/TF1, andπ:E→M1the bundle projection. Identify M1with the image of the zero section of E. Fixing any Riemannian metric on M, there are induced Riemannian metrics on the Mi, and an induced Riemannian structure on E. For eachr >0, letBr⊂Edenote the open disk bundle of radius roverM1. Then there is an R >0 and an open neighborhood Uof the support of finM1such that, ifV=π−1(U)∩BR, the restriction of ι1toUcan be extended to a map of foliated manifolds, ˜ ι1: (V,π∗F1|V)→(M,F), which is defined over each x∈M1 as a composite of the restriction of the canonical homomorphism (ι∗ 1TF/TF1)x→Tι1(x)F/ι1∗TxF1≡(ι1∗TxF1)⊥∩TxF, and the exponential map of the leaves of Fdefined on the ball of radius Rcentered at zero in (ι1∗TxF1)⊥∩TxF. By elementary properties of the exponential map and sinceι1is transversely regular with codim F1= codim F,Rcan be chosen so that ˜ι1is a local diffeomorphism and ˜ ι∗ 1F=π∗F1|V. Eis of rankk, and with an induced orientation. The representatives Φ nof its Thom class, given by Lemma 4.1, can be assumed to be supported in BR. The Φ n are of degree k, closed and compactly supported in the vertical direction, i.e. with compactly supported restrictions to the fibers. Moreover all the Φn|BRare pairwise cohomologousin the complex of forms in Ω·(BR) which are compactly supported in the vertical direction. On the other hand, fis basic and compactly supported. So theπ∗fΦnrestrict to leafwise closed forms αn∈Ωk c(π∗F1|V) which are pairwise cohomologous. Thus the ˜ ι1∗αn∈Ωk c(F) are leafwise closed and define the same classζ∈ Hk c(F). LetU1,...,U m, be an open cover of the support of finUsuch that each ι1:Uj→Mis an embedding, j= 1,...,m. The above Rcan be chosen small enough so that the ˜ ι1:Vj=π−1(Uj)∩BR→˜ι1(Vj) are diffeomorphisms. Take a decomposition f=f1+···+fmwith eachfjcompactly supported in Uj, and let αn,j∈Ωk c(π∗F1|Vj) be the restriction to the leaves of π∗fjΦn. Then, by definition, ˜ι1∗αn=βn,1+···+βn,m,14 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR where each βn,j∈Ωk c(F) is the extension by zero of the forms in Ωk c(F|˜ι1(Vj)) which correspond to the αn,j|Vjby ˜ι1. Givenι2: (M2,F2)→(M,F), we use the same notation as in the preamble of Definition 2.1. We can clearly assume the Ujare contained in N1. Letι=ι1σ1= ι2σ2:T→M. There is a canonical isomorphism ι∗TF∼=σ∗ 1TF1⊕σ∗ 2TF2 because the ιi|Niintersect transversely in F. Soσ∗ 1E∼=σ∗ 2TF2canonically. This isomorphism will be considered as an identity. LetπT:σ∗ 1E→Tbe the pull-back vector bundle projection, and ˜ σ1:σ∗ 1E→E the canonical homomorphism. Identify Tto the image of the zero section of σ∗ 1E. For eachj, take a relatively compact open subset Oj⊂σ−1 1(Uj) containing the compact set supp σ∗ 1fj∩suppDT. The above Rcan be chosen small enough so that σ2:Oj→N2hasan extensionto alocaldiffeomorphism ˜ σ2:π−1 T(Oj)∩˜σ−1 1(BR)→ N2defined as the composite of the restriction of the canonical homom orphism σ∗ 1E≡σ∗ 2TF2→TF2, and the exponential map of the leaves of F2defined on the tubular neighborhood of radius Rof the zero section in TF2. In this way, ˜ σ2maps each fiber of πTinto a leaf of F2. Observe that the diagram ˜σ−1 1(BR)∩π−1 T(Oj)˜σ1− −−− →Vj ˜σ2/arrowbt/arrowbt˜ι1 N2ι2− −−− →M is obviously non-commutative in general. This is the main technical diffic ulty. To solve it, we have chosen the Φ nso that their supports concentrate around M1and satisfy the needed properties at the limit (Lemma 4.1 and Corollary 4.2 ). We need the observation that (12) σ2σ−1 1(A) = (ι2|N2)−1ι1(A) for any subset A⊂N1, as can be easily checked. Using the compactness of BR∩π−1(suppfj) and since suppfj=/intersectiondisplay 0<r<RBr∩π−1(suppfj), we easily get ι1(suppfj) =/intersectiondisplay 0<r<R˜ι1/parenleftbig Br∩π−1(suppfj)/parenrightbig . Therefore/intersectiondisplay 0<r<RsuppD∩ι−1 2˜ι1/parenleftbig Br∩π−1(suppfj)/parenrightbig = suppD∩ι−1 2ι1(suppfj) = suppD∩σ2σ−1 1(suppfj) =σ2(σ−1 2(suppD)∩σ−1 1(suppfj)) =σ2(suppDT∩suppσ∗ 1fj), where the second equality follows by (12). Then, since the suppD∩ι−1 2˜ι1/parenleftbig Br∩π−1(suppfj)/parenrightbig are compact, and since ˜σ2/parenleftbig ˜σ−1 1(BR)∩π−1 T(Oj)/parenrightbigLEAFWISE REDUCED COHOMOLOGY 15 is an open neighborhood of σ2(suppDT∩suppσ∗ 1fj), there is an r<Rsuch that suppD∩ι−1 2˜ι1/parenleftbig Br∩π−1(suppfj)/parenrightbig ⊂˜σ2/parenleftbig ˜σ−1 1(BR)∩π−1 T(Oj)/parenrightbig . So suppD∩suppι∗ 2βn,j⊂˜σ2(Wj) for large enough n, where Wj= ˜σ−1 2ι−1 2˜ι1(Vj)∩˜σ−1 1(BR)∩π−1 T(Oj). We can assume this holds for every n. Hence there is some ωn,j∈Ωk c(F2) which is supported in ˜ σ(Wj) and has the same restriction to some neighborhood of supp D asι∗ 2βn,j. IfFπTis the foliation on σ∗ 1Edefined by the fibers of πT, there is some γn,j∈Ωk c(FπT|Wj) such that (˜ σ2|Wj)∗ωn,j=ι∗ 2βn,j. We get (13) D(ι∗ 2ζ) =/summationdisplay jDT/parenleftBigg/integraldisplay πT|Wjγn,j/parenrightBigg by definition. Let hj:Wj→Ebe the immersion given by the composite Wj˜σ2− −−− →ι−1 2˜ι1(Vj)ι2− −−− →˜ι1(Vj)˜ι−1 1− −−− →Vj⊂E . Clearlyhjis an extension of σ1:Oj→N1⊂M1⊂E, andγn,j=h∗ j(π∗fjΦn) aroundWj∩π−1 T(σ−1 1(suppfj)∩suppDT). Moreover the homomorphism between the normal bundles of OjandM1, defined by the differential of hj, restricts to isomorphismsonthefibers. Theseisomorphismsareorientationpre servingonfibers over points with ε= 1, and orientation reversing on fibers over points with ε=−1. Therefore/integraltext πT|Wjγn,jconverges to εσ∗ 1fjonσ−1 1(suppfj)∩suppDTwith respect to theC∞topology by Corollary 4.2. Hence (13) is equal to /a\}b∇acketle{t(ι1,f),(ι2,D)/a\}b∇acket∇i}ht, and the proof is complete. Remark3.Observe that, in Proposition 3.1, ζhas representativessupported in any neighborhood of ι1(M1). Thus, in Theorem 2.2, the linearly independent classes ζm∈ H· c(F) also have representatives supported in any neighborhood of ι1(M1). 5.Case where the leaves have homology classes with non-trivia l intersection This section will be devoted to the proof of Corollary 2.3. LetMbetheambientmanifoldof F. Letπi:ι∗ i(TM/TF)→Kidenotethepull- backvectorbundle projection, andidentify Kitothe imageofitszerosection. Fix a Riemannian metric on Mand, for some R>0, letMibe the tubular neighborhood of radiusRaroundKiinι∗ i(TM/TF). SuchRcan be chosen so that the maps ˜ιi:Mi→Mare well defined as composites of the restrictions of the canonical homomorphisms ι∗ i(TM/TF)→TM/TF ≡(TF)⊥, and the restriction of the exponential map of Mto the tubular neighborhood of radius Rof the zero section ofTF⊥. ChooseRsmall enough so that ˜ ιi:π−1 i(xi)∩Mi→˜ιi(π−1 i(xi)∩Mi) is an embedded transversal of Ffor eachiand eachxi∈Mi. Observe that ˜ι1(π−1 1(x1)∩M1) = ˜ι2(π−1 2(x2)∩M2) ifι1(x1) =ι2(x2). The ˜ιiare thus transverse toF, and the Fi= ˜ι∗ iFhave the same codimension as F. Then ˜ιiare transversely16 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR regular immersions of foliated manifolds. By deforming the ιiif needed, we can supposethe ιiintersecteachothertransversely,andthusthe˜ ιiintersecttransversely inF. Moreover the orientations of the Kiinduce orientations of the Fi. EachKiis a closed leaf of Fiwhose holonomy representation is given by the composite (3). So the holonomy group of Kiis given by germs of local isometries. HenceFiisaRiemannianfoliationaround Ki, asfollowseasilyfrom[14, Theorem2 in Chapter IV]. (See also [12, Theorem 2.29 in Chapter II] or [17].) We ca n assume the whole Fiis Riemannian, which can thus be described as follows [16, 26]. Fix a transverse Riemannian structure of Fi. LetQibe theO(q)-principal bundle overMiof transverse orthonormal frames of Fiwith the transverse Levi-Civita connection, and ˆFithe horizontal lifting of FitoQi[24, 25]. Let Pibe a leaf closure of ˆFioverKi. ThenPiis anHi-principal bundle over Kifor some closed subgroupHi⊂O(q). For the open disk B⊂Rqof radiusRcentered at the origin, we can assume Mi≡(Pi×B)/Hias fiber bundles over Ki, where the Hi-action on P2×Bis the diagonal one; i.e. ( z,v)h= (zh,h−1v) for (z,h)∈Pi×Bandh∈Hi. Moreover the above identity can be chosen so that Fiis identified to the foliation whose leaves are the projections of products of leaves of ˆFiinPiand points in B. (This description is simpler than the one in [16] and [26] because the le af closure Kiis just a compact leaf.) Consider the transverse Riemannian structure of each Fidefined by the Eu- clidean metric on Busing the above description. Since the elements in the image of the composites (3) are germs of local isometries for the same me tric onRq, the composite B≡π−1 2(x2)∩M2˜ι2−→˜ι2(π−1 2(x2)∩M2) = ˜ι1(π−1 1(x1)∩M1)˜ι−1 1−→π−1 1(x1)∩M1≡B is an isometry around the origin for all ( x1,x2)∈K1×K2withι1(x1) =ι2(x2). We can assume such composite is an isometry on the whole B, which will be denoted byφx2,x1. With the above description, any compactly supported basic functio nfofF1can be canonically considered as an H1-invariant compactly supported function on B, and any compactly supported holonomy invariant transverse distr ibutionDofF2 can be canonically considered as a compactly supported H2-invariant distribution onB. For suchfandD, we clearly have (14) /a\}b∇acketle{t(˜ι1,f),(˜ι2,D)/a\}b∇acket∇i}ht=/summationdisplay ε(x1,x2)D(φ∗ x2,x1f), where the sum runs over the pairs ( x1,x2)∈K1×K2withι1(x1) =ι2(x2). Here ε(x1,x2) =±1 depending on whether the identity Tιi(xi)B≡ι1∗Tx1K1⊕ι2∗Tx2K2 isorientationpreservingororientationreversing. Let fmbeasequenceofcompactly supportedO(q)-invariant functions in Bwith integral equal to one and pairwise disjoint supports, and let µmbe the restriction of the Euclidean measure to the support offm. Then /a\}b∇acketle{t(˜ι1,fm),(˜ι2,µn)/a\}b∇acket∇i}ht=/a\}b∇acketle{tι1,ι2/a\}b∇acket∇i}ht/integraldisplay Bfmdµn by (14), where /a\}b∇acketle{tι1,ι2/a\}b∇acket∇i}htis theintersectionnumberof ι1andι2inB. Sodim Hk2 c(F) = ∞by Theorem 2.2. Similarly, dim Hk1 c(F) =∞, which completes the proof.LEAFWISE REDUCED COHOMOLOGY 17 6.Case of suspension foliations Proof of Corollary 2.5. Recall the notation used for suspension foliations in the statement of Corollary 2.5, and consider the fiber bundles Mi=ι∗ iMoverKi. Each canonical map ˜ ιi:Mi→Mis transverse to F, and let Fi= ˜ι∗ iF. Then ˜ιiare transverselyregularimmersionsoffoliatedmanifolds. Bydeforming theιiifneeded, we can suppose the ιiintersect each other transversely, thus the ˜ ιiintersect each other transversely in F. Moreover the orientations of the Kiinduce orientations of theFi. The group of deck transformations of each pull-back covering map ι∗ iL→Kiis isomorphic to Γ i, andFiis canonically isomorphic to the corresponding suspension foliation given by the restriction to Γ iof the Γ-action on T. Hence the fmcan be canonically considered as compactly supported basic functions of F1, and theDn can be canonically considered as holonomy invariant transverse dist ributions of F2. The ˜ι2: suppDn→Mare clearly proper, and we easily get /a\}b∇acketle{t(˜ι1,fm),(˜ι2,Dn)/a\}b∇acket∇i}ht=/a\}b∇acketle{tι1,ι2/a\}b∇acket∇i}htDn(fm). Therefore the result follows from Theorem 2.2. /square Proof of Corollary 2.6. LetAnbe a sequence of Γ 1-saturated open subsets of T with non-trivial µ-measure and pairwise disjoint Γ 2-saturations. Clearly, there are open sets BnofTwith positive µ-measure and such that Bn⊂An. Since the closure of Γ 1in the group of diffeomorphisms of Tis a compact Lie group, there exists a sequence of non-negative Γ 1-invariant functions fnonTsuch that Bn⊂suppfn⊂An. Letµnbe the Γ 2-invariant measure on Tdefined as the product of µand the characteristic function of the closure of the Γ 2-saturation of suppfn. Then/integraltext Tfmdµn/\e}atio\slash= 0 if and only if m=n, and the result follows by Corollary 2.5. /square Proof of Corollary 2.7. Since some lift of ctoLjoins two distinct points of its end set,Lis disconnected by some codimension one immersed closed submanifold , ι:K→L, such that candπιdefine homology classes of Bwith non-trivial intersection. Clearly, the composite π1(K)π1(πι)− −−− →π1(B)h− −−− →Γ is trivial, and the image of the composite π1(S1)π1(c)− −−− →π1(B)h− −−− →Γ is/a\}b∇acketle{ta/a\}b∇acket∇i}ht. Take a sequence Anof disjoint non-trivial H-invariant open subsets of T. SinceHis an abelian compact Lie group (a torus), there is an H-invariant prob- abilistic measure supported in any H-orbit inT. Take thus one of such measures µnsupported in each An. Then the result follows from Corollary 2.6 by taking as µthe sum of the µn. /square To prove Proposition 2.8, we use the following. Lemma 6.1. LetΓbe a finitely generated group, and Xa connected T1topological space. For any continuous action of ΓonX, a finite union of orbits is dense if and only if so is each orbit in the union.18 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR Proof.Takex1,...,x n∈Xsuch that X=Γx1∪...∪Γxn=Γx1∪...∪Γxn. Each orbit closure Γxican be decomposed as a disjoint union of sets Li=/intersectiondisplay F(Γ\F)xi, Ii=Γxi\Li, whereFruns over the finite subsets of Γ. We have X=L∪I, whereL=/uniontextn i=1Li andI=/uniontextn i=1Ii. Moreover, since Lis saturated we have L∩I=∅. SoI=∅ becauseXisT1and connected. (If we had I/\e}atio\slash=∅, for anyy∈I,{y}would be closed inXbecauseXisT1. But since Lis closed and I=X\Lis discrete, {y} would be also open in X. ThusXwould not be connected.) Therefore X=Land Li=Γxifor eachi. But each Liis closed in X, andLi∩Lj/\e}atio\slash=∅impliesLi=Lj, obtainingX=Lifor everyiby the connectedness of X. /square Proof of Proposition 2.8. Clearly,ifthe C-orbitsaredensein X,soaretheΓ-orbits. Reciprocally, suppose the Γ-orbits are dense. By a theorem of Sta llings [31], there is a finite normal subgroup F⊂Γ such that Γ 1= Γ/Fis isomorphic either toZor to the diedric group Z2∗Z2. The action of Γ on Xdefines an action of Γ1on the connected T1spaceX1=X/Fwith dense orbits. Since Cis infinite, so is its projection C1to Γ1, and any infinite subgroup of such Γ 1is of finite index. ThereforeanyΓ 1-orbitinX1isafiniteunionof C1-orbits,andthusthe C1-orbitsare dense inX1by Lemma 6.1. This implies the density of the CF-orbits inXbecause the canonical projection of XontoX1is open and continuous. But any CF-orbit is a finite union of C-orbits. Hence the C-orbits are dense by Lemma 6.1. /square 7.Case of Lie foliations with compact semisimple structural L ie algebra Theorem 2.9 will be proved in this section (Corollaries 7.16 and 7.17). 7.1.Construction of a spectral sequence for an arbitrary Lie fol iation on a closed manifold. LetFbe a Lie foliationwith dense leaveson a closedmanifold M. Letgbe the structural Lie algebra of F, andGthe simply connected Lie group with Lie algebra g. Letπ:/tildewiderM→Mbe the universal covering map. Then the leaves of/tildewideF=π∗Fare the fibers of a fiber bundle D:/tildewiderM→G. It will be convenient to consider the right action of π1(M) on/tildewiderMby deck transformations and the left action ofGon itself by left translations. Thus Dis anti-equivariant with respect to the holonomy homomorphism h:π1(M)→G; i.e.,D(˜xγ) =h(γ)−1D(˜x) [11]. The density ofthe leavesimplies the density of Γ = h(π1(M)) inG. The homomorphism hdefines an action of π1(M) onGby left translations, yielding the corresponding suspension foliation GonN=/parenleftBig /tildewiderM×G/parenrightBig /π1(M) (defined as in Section 6). Gis a Lie foliation with the same transverse structure as F, given by ( G,Γ). The section (id ,D) :/tildewiderM→/tildewiderM×Gisπ1(M)-equivariant: (id,D)(˜xγ) = (˜xγ,D(˜xγ)) = (˜xγ,γ−1D(˜x)) = (˜x,D(˜x))γ . Thus (id,D) defines a section s:M→N, andNis trivial as principal G-bundle overM. Clearlysis transverse to G, ands∗G=F.LEAFWISE REDUCED COHOMOLOGY 19 LetD:/tildewiderM×G→Gbe defined by D(˜x,g) =g−1D(˜x). SuchDisπ1(M)- invariant: D((˜x,g)a) =D(˜xa,h(a)−1g) =g−1h(a)D(˜xa) =g−1D(˜x). SoDdefines a map DN:N→G. ClearlyDNs= const e, whereeis the identity element inG. Moreover DNisG-anti-equivariant: DN([˜x,g]g′) =DN([˜x,gg′]) = (gg′)−1D(˜x) = (g′)−1DN([˜x,g]). ThereforeDNis the composite of the second factor projection of the trivializatio n ofN→Mdefined bysand the inversion map on G. Let/tildewideFalso denote the foliation on Ndefined by the lifting of Fto all the leaves ofG./tildewideFis a subfoliation of Gwhose leaves are the intersections of the leaves of G with all the translations of s(M). Letν⊂TGbe aG-invariant subbundle so that TG=ν⊕T/tildewideF. We get /logicalanddisplay TG∗=/logicalanddisplay ν∗⊗/logicalanddisplay T/tildewideF∗, and thus there is a bigrading of Ω = Ω( G) defined by Ωu,v=C∞/parenleftBiggu/logicalanddisplay ν∗⊗v/logicalanddisplay T/tildewideF∗/parenrightBigg , u,v∈Z. For simplicity, dGwill be denoted by d. There is a decomposition of das sum of bihomogeneous components, d=d0,1+d1,0+d2,−1, where each double subindex denotes the corresponding bidegree. From d2= 0 we get (15) d2 0,1=d2 2,−1=d0,1d1,0+d1,0d0,1= 0, (16) d1,0d2,−1+d2,−1d1,0=d2 1,0+d0,1d2,−1+d2,−1d0,1= 0. The decreasing filtration of (Ω ,d) by the differential ideals (17) Fl= Ωl,·∧Ω, depends only on/parenleftBig G,/tildewideF/parenrightBig ; it could be defined without using ν. So we get a spectral sequence (Ei,di) which converges to H·(G). As for the spectral sequence of a foliation (see e.g. [1]), in this case there are canonical identities (18) ( E0,d0)≡(Ω,d0,1),(E1,d1)≡(H(Ω,d0,1),d1,0∗). TheC∞topology on the space of differential forms induces a topology on ea chEi which is not Hausdorff in general. At eachz∈Nwe have DN∗:νz∼=− −−− →TDN(z)G. So for each X∈gthere is a well defined vector field Xν∈C∞(ν) which isDN- projectable and such that DN∗Xν=X. SuchXνisG-invariant since Xν zg∈νzg and DN∗(Xν zg) =g−1DN∗Xν z=g−1XDN(z)=Xg−1DN(z)=XDN(zg). LetθXandiXrespectively denote the Lie derivative and interior product on Ω with respect to Xν. (We are considering θXandiXas operators on the leaves of G, but preserving smoothness on N.) By comparing bidegrees in the usual formulas20 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR that relate Lie derivatives, interior products and the de Rham deriv ative, we easily get d0,1iX+iXd0,1= 0, (θX)0,0d0,1=d0,1(θX)0,0, i[X,Y]= (θX)0,0iY−iY(θX)0,0, (θX)0,0=d1,0iX+iXd1,0, (θ[X,Y])0,0= (θXθY−θYθX)0,0−d0,1iΞ(X∧Y)−iΞ(X∧Y)d0,1, where Ξ :/logicalandtext2g→C∞/parenleftBig T/tildewideF/parenrightBig is given by Ξ(X∧Y) = [Xν,Yν]−[X,Y]ν. Therefore we get the operation ( g,i1,θ1,E1,d1), wherei1X≡iX∗andθ1X≡ (θX)0,0∗according to (18), and the algebraic connection D∗ N:g∗→E1,0 1⊂Ω1,0 [13]. Then Eu,v 2∼=Hu(g;θ1:g→End(E0,v 1)). Letφ:N×g→Nbe defined by φ(z,X) =Xν 1(z), whereXν tdenotes the uni- parametric group of transformations defined by Xν, considered as group of trans- formations of the leaves of Gpreserving smoothness on N. Then the following diagram is commutative N×gφ− −−− →N DN×exp/arrowbt/arrowbtDN G×G− −−− →G, where the lowest map denotes the operation on G. (This follows because Xt= Rexp(tX)for allX∈g.) 7.2.Tensor product decomposition of E2whengis compact semisimple. From now on suppose gis compact semisimple, and thus Gis compact [28]. Theorem 7.1. With the above notations, Eu,v 2∼=Hu(g)⊗E0,v 2=Hu(g)⊗(E0,v 1)θ1=0. The resultfollowswith the sametypeofargumentsasin thosegivenin Sections2 and 3 of [2] to prove Theorem 3.5 in [2]. We will indicate the main steps in th e proof because some of them will be needed later. Consider the canonical biinvariant metric on G[28, Chapter 6], and let C⊂G andC∗⊂gbe the cut locus and tangential cut locus corresponding to the iden tity elemente∈G. LetB∗be the radial domain in gbounded by C∗, and letB= exp(B∗). From the general properties of the cut locus we have C=∂B=G\B, exp :B∗→Bis a diffeomorphism, CandC∗have Lebesgue measure zero, and B∗ is compact (since so is G) [22, 21]. Consider the compact space F={(X,Y,Z)∈B∗3: exp(X) exp(Y) = exp(Z)} ⊂g3, and for each X∈B∗the compact slice FX={(Y,Z)∈g2: (X,Y,Z)∈F} ⊂g2. Smoothness on FandFXwill refer to the smoothness obtained by considering these spaces as subspaces of g3andg2, respectively.LEAFWISE REDUCED COHOMOLOGY 21 Letι:g2→g2be the involution ( Y,Z)/mapsto→(Z,Y). Fora= exp(X) we also have the smooth map JX:B∩L−1 aB→FXgiven by JX(g) = (log(g),log(ag)), where log = exp−1:B→B∗. LetWX=JX(B∩L−1 aB)⊂FX. Lemma 7.2 ([2, Proposition 2.2]) .We have: (i)WXis open inFXandJX:B∩L−1 aB→WXis a diffeomorphism. (ii)ι(FX) =F−X, and the diagram B∩L−1 aBJX− −−− →FX La/arrowbt/arrowbtι B∩LaBJ−X− −−− →F−X is commutative. ForX,Y∈B∗letWX,Y=JX(B∩L−1 aB∩L−1 bB)⊂FX, wherea= exp(X) andb= exp(Y). We have the diffeomorphism JX,Y=JYJ−1 X:WX,Y→WY,X. Let ∆ be the unique biinvariant volume form on Gsuch that/integraltext G∆ = 1, which defines a Haar measure µonG. Then for each X∈B∗letµXbe the Borel measure onFX, concentrated on WX, where it corresponds to µbyJX. Corollary 7.3 ([2, Proposition 2.3]) .We have: (i)µX(FX) =µX(WX) =µX(WX,Y) =µ(B∩L−1 aB∩L−1 bB) =µ(B∩L−1 aB) =µ(G) = 1 (ii)µXcorresponds to µ−Xbyι:FX→F−X. (iii)µXcorresponds to µYbyJX,Y:WX,Y→WY,X. LetI= [0,1], and define continuous maps σ,η:F×I→Gby setting σ(ξ,t) = exp(tZ), η(ξ,t) =/braceleftBigg exp(2tX) if t∈I1= [0,1/2] exp(X) exp((2t−1)Y) ift∈I2= [1/2,1], whereξ= (X,Y,Z)∈F. The map σis smooth, and so are the restrictions of ηto eachF×Ii(i= 1,2). Lemma 7.4 ([2, page 178]) .There is a finite open cover Q1,...,Q kofF, and continuous maps Hj:Qj×I×I→Gwith smooth restrictions to each Qj×Ii×I, i= 1,2,j= 1,...,k, so that Hj(·,·,0) =σ|Qj×I,Hj(·,·,1) =η|Qj×I, Hj(ξ,0,s) =efor alls∈Iandξ∈Qj, Hj(ξ,1,s) = exp(Z)for alls∈Iandξ= (X,Y,Z)∈Qj. Lemma 7.5. For eachj= 1,...,kthere exists a unique continuous map Hj:N×Qj×I×I→N with smooth restrictions to each N×Qj×Ii×I,i= 1,2, such that (i)DNHj(z,ξ,t,s) =DN(z)Hj(ξ,t,s), (ii)Hj(z,ξ,0,s) =z, (iii) (d/dt)Hj(z,ξ,t,s)∈νfort/\e}atio\slash= 1/2. Moreover for ξ= (X,Y,Z)∈Qjwe have22 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR (iv)Hj(·,ξ,1,0) =φZ, (v)Hj(·,ξ,1,1) =φYφX (vi)Hj(z,ξ,1,s)∈D−1 N(D(z) exp(Z))for allz∈Nand alls∈I. Proof.It is completely similar to the proofs of Lemmas 3.1 and 3.2 in [2]. /square Therefore, for all ξ= (X,Y,Z)∈Qj,Hj(·,ξ,1,·) :N×I→Nis an/tildewideF-integrable homotopy of φZtoφYφX[10]. Hence the corresponding homotopy operator in Ω preservesthefiltration, andthusits(0 ,−1)-bihomogeneouscomponent kj,ξ: Ω→Ω satisfies (φ∗ Xφ∗ Y−φ∗ Z)0,0=d0,1kj,ξ+kj,ξd0,1. Define the operators ρ,λ: Ω→Ω by setting ρ(α) =/integraldisplay B∗φ∗ Xα∆∗(X), λ(α) =/integraldisplay B∗ΦXα∆∗(X), where ∆∗= exp∗∆ and Φ Xis the homogeneous operator of degree −1 on Ω associ- atedtothehomotopy φtX(t∈I) [9]. Theoperators ρandλarelinearhomogeneous of degrees 0 and −1, respectively, satisfying ρ−id =dλ+λd. Moreover, since φtX preserves the pair of foliations/parenleftBig G,/tildewideF/parenrightBig (becauseXνis an infinitesimal transforma- tion of/parenleftBig G,/tildewideF/parenrightBig ), ΦXreduces the filtration at most by a unit. Therefore the bihomo- geneous operators ρ1≡ρ0,0∗andλ1≡λ−1,0∗onE1satisfyρ1−id =d1λ1+λ1d1. Forα∈Ω andX∈B∗, by Lemma 7.2 and Corollary 7.3 we have φ∗ Xρ(α) =/integraldisplay FXφ∗ Xφ∗ YαdµX(Y,Z), ρ(α) =/integraldisplay WX,−Xφ∗ YαdµX(Y,Z) =/integraldisplay W−X,Xφ∗ Yαdµ−X(Y,Z) =/integraldisplay F−Xφ∗ Yαdµ−X(Y,Z) =/integraldisplay FXφ∗ ZαdµX(Y,Z). So (19) ( φ∗ Xρ−ρ)α=/integraldisplay FX(φ∗ Xφ∗ Y−φ∗ Z)αdµX(Y,Z). Take a smooth partition of unity f1,...,f kofFsubordinated to the open cover Q1,...,Q k. Then the fj(X,·,·) form a partition of unity of FXsubordinated to the open cover given by the slices Qj,X={(Y,Z)∈g2: (X,Y,Z)∈Qj}. Let ΨX: Ω→Ω be the (0 ,−1)-bihomogeneous linear operator given by ΨXα=k/summationdisplay j=1/integraldisplay Qj,Xkj,ξαfj(ξ)dµX(Y,Z), whereξ= (X,Y,Z) for each (Y,Z)∈Qj,X. From (19) we get (20) ( φ∗ Xρ−ρ)0,0=d0,1ΨX+ΨXd0,1.LEAFWISE REDUCED COHOMOLOGY 23 Lemma 7.6. ΨXαdepends continuously on X∈B∗for eachα∈Ωfixed. Proof.It is completely analogous to the proof of Lemma 3.3 in [2]. /square Lemma 7.7. Forα∈Ω,X∈gandt∈Rwe have φ∗ tXα=α+/integraldisplayt 0φ∗ sXθXαds=α+θX/integraldisplayt 0φ∗ sXαds. Proof.It is completely analogous to the proof of Lemma 3.4 in [2]. /square Lemma 7.8. ρ1(E1) = (E1)θ1=0, and ρ1∗:E2∼=− −−− →H((E1)θ1=0). Proof.First, we shall prove that ρ1(E1)⊂(E1)θ1=0. Take any α∈ker(d0,1) defining [α]∈E1. If [α]∈ρ1(E1), we can suppose α=ρ0,0βfor someβ∈ker(d0,1). Then (21) ( φ∗ X)0,0α−α=d0,1ΨXβfor allX∈B∗ by (20). Thus Lemmas 7.6 and 7.7 yield (θX)0,0α=d0,1/parenleftbigg ΨXβ−(θX)0,0/integraldisplay1 0ΨsXβds/parenrightbigg as in [2, page 181]. Therefore ρ1([α])∈(E1)θ1=0. Letι: (E1)θ1=0→E1be the inclusion map. If [ α]∈(E1)θ1=0, since (θX)0,0 depends linearly on X∈g, there is a linear map X/mapsto→βXofgto Ω so that (θX)0,0α=d0,1βXfor allX∈g. Thus by Lemma 7.7 we get ρ0,0α=α+d0,1/integraldisplay B∗/integraldisplay1 0(φ∗ sX)0,0βXds∆∗(X), yieldingρ1ι= id. In particular ρ1(E1) = (E1)θ1=0. We also have ιρ1−id = d1λ1+λ1d1, and the result follows. /square End of the proof of Theorem 7.1. SinceGiscompact,therepresentation θgissemisim- ple [13, Sections 4.4 and 5.12]. So H((E1)θ1=0)∼=H(g)⊗(E0,· 1)θ1=0 by [13, Theorem V in Section 4.11, and Section 5.26]. The result now follo ws from Lemma 7.8. /square 7.3.Relation between H·(F)andE2. Theorem 7.9. With the above notations, H·(F)∼=E0,· 2. To begin with the proof of Theorem 7.9, the section s:M→Ndefines a homomorphism ( s∗)1:E0,· 1→H·(F) sinces∗d0,1=dFs∗. By restricting ( s∗)1, we get (s∗)2:E0,· 2= (E0,· 1)θ1=0→H·(F). We will prove that ( s∗)2is an isomorphism. For anyX∈gsetsX=φXs:M→N, which is an embedding, but not a section ofπNin general. Nevertheless sX(M) =s(M) exp(X). Analogously to s, the mapsXalso defines ( s∗ X)1:E0,· 1→H·(F). LetUXbe the neighborhood of sX(M) given by UX=/uniondisplay Y∈B∗φYsX(M) =sX(M)B=D−1 N(exp(X)B).24 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR For eachX∈gand eachx∈M,sXdefines an isomorphism sX∗:TxF∼=− −−− →TsX(x)/tildewideF. So s∗ X:/logicalandtextTsX(x)/tildewideF∗ ∼=− −−− →/logicalandtextTxF∗. For eachω∈Ω(F), letωXbe the unique smooth section of/logicalandtextT/tildewideF∗oversX(M) such thats∗ XωX=ω. DefineTXω∈Ω0,·(G|UX) by setting (TXω)(φYsX(x)) = (φ∗ −Y)0,0ωX(sX(x)) forY∈B∗andx∈M. This is well defined since ( x,Y)/mapsto→φYsX(x) is a dif- feomorphism of M×B∗ontoUX. Moreover d0,1TX=TXdFsinced0,1≡d/tildewideFon Ω0,·≡Ω·/parenleftBig /tildewideF/parenrightBig , and (φYsX)∗/tildewideF=Ffor allX,Y∈g. Therefore TXdefines a map TX∗:H·(F)→E0,· 1(G|UX). The inclusion map ιX:UX→Ninduces (ι∗ X)1:E0,· 1→E0,· 1(G|UX). Lemma 7.10. Ifζ∈(E0,· 1)θ1=0, thenTX∗(s∗ X)1ζ= (ι∗ X)1ζ. Proof.By Lemma 7.8 we have ρ1ζ=ζ. We thus can choose forms α,γ∈Ω0,·such thatd0,1α= 0,ζ= [α], andα=ρ0,0α+d0,1γ. Then (21) yields (φ∗ Y)0,0(α−d0,1γ)−(α−d0,1γ) =d0,1ΨYα for anyY∈B∗. So (22) ( φ∗ Y)0,0α−α=d0,1(ΨYα+(φ∗ Y)0,0γ−γ). Clearly (s∗ Xα)X=α|sX(M). Hence (TXs∗ Xα)(φYsX(x)) = (φ∗ −Y)0,0(α(sX(x))) = (α+d0,1(Ψ−Yα+(φ∗ −Y)0,0γ−γ))(φYsX(x)) by (22). But since each φYsX(M) is/tildewideF-saturated, d0,1≡d/tildewideFcommutes with the restriction to each φYsX(M). Therefore we get TXs∗ Xα=α+d0,1ηX onUX, whereηXis the (0,·)-form onUXdefined by ηX(φYsX(x)) = (Ψ −Yα+(φ∗ −Y)0,0γ−γ)(φYsX(x)), which finishes the proof. /square SinceGis compact, there is a finite sequence 0 = X1,X2,...,X lof elements of gsuch that G=B∪exp(X2)B∪···∪exp(Xl)B . LetUj=UXjTj=TXj,sj=sXjandιj=ιXjforj= 1,...,l. ThenN= U1∪···∪Ul. Leth1,...,h lbe a smooth partition of unity of Gsubordinated to the open cover exp( X1)B,...,exp(Xl)Bso thath1(e) = 1. Then D∗ Nh1,...,D∗ Nhlis a partition of unity of N subordinated to U1,...,U l. Forω∈Ω(F), defineTω∈Ω0,·by setting Tω=l/summationdisplay j=1D∗ NhjTjω.LEAFWISE REDUCED COHOMOLOGY 25 Since each D∗ Nhjis constant along the leaves of /tildewideF, we getd0,1T=TdF.SoT defines a map T∗:H·(F)→E0,· 1. Lemma 7.11. Ifζ∈(E0,· 1)θ1=0, thenT∗(s∗)1ζ=ζ. Proof.For eachX∈g, let (φ∗ X)1:E1→E1be the homomorphism defined by φ∗ X ((φ∗ X)1≡(φ∗ X)0,0∗). SincesX=φXs, by (21) we have (s∗ X)1ζ=s∗ 1(φ∗ X)1ζ=s∗ 1ζ . Therefore, by Lemma 7.10, (ι∗ j)1ζ=Tj∗(s∗ j)1ζ=Tj∗(s∗)1ζ forj= 1,...,l. So, ifζ= [α] forα∈Ω0,·withd0,1α= 0, there is some βj∈Ω0,· for eachjsuch thatα−Tjs∗α=d0,1βjoverUj. Let β=l/summationdisplay j=1D∗ Nhjβj∈Ω0,·. Since eachD∗ Nhjis constant on the leaves of /tildewideFandd0,1≡d/tildewideF, we get d0,1β=l/summationdisplay j=1D∗ Nhjd0,1βj =l/summationdisplay j=1D∗ Nhj(α−Tjs∗α) =α−Ts∗α, and the proof is complete. /square Lemma 7.12. (s∗)2:E0,· 2→H·(F)is surjective. Proof.Take anyω∈Ω(F) withdFω= 0, and take any function f≥0 compactly supported in Bsuch that/integraltext Bf(g)∆(g) = 1. Then α=D∗ NfT1ωis a (0,·)-form compactly supported in U1and satisfying d0,1α= 0. Soαdefines a class ζ∈E0,· 1. We shall prove that ( s∗)1ρ1ζ= [ω]. Forx∈MandY∈B∗we have α(φYs(x)) =f(exp(Y))(φ∗ −Y)0,0(ωX1(s(x))). So ((φ∗ Y)0,0α)(s(x)) =f(exp(Y))ωX1(s(x)), yielding (ρ0,0α)(s(x)) =/integraldisplay B∗((φ∗ Y)0,0α)(s(x))∆∗(Y) =ωX1(s(x))/integraldisplay B∗f(exp(Y))∆∗(Y) =ωX1(s(x))/integraldisplay Gf(g)∆(g) =ωX1(s(x)). Therefores∗ρ0,0α=s∗ωX1=ω, and the proof follows. /square Corollary 7.13. T∗(H·(F))⊂E0,· 2.26 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR Proof.It follows directly from Lemmas 7.11 and 7.12. /square End of the proof of Theorem 7.9. By Corollary 7.13 we can consider T∗:H·(F)→ E0,· 2. By Lemma 7.11 we have T∗(s∗)2= id. On the other hand, ( s∗)2T∗= id because (D∗ Nh1)(s(x)) = 1 for all x∈Msinceh1(e) = 1. So ( s∗)2is an isomor- phism. /square Corollary 7.14. H1(F)∼=H1(G)andH1(F)∼=H1(G). Proof.Theorem 7.1 yields E2,0 2∼=H2(g)⊗E0,0 2= 0 since gis compact semisimple. SoE0,1 2=E0,1 ∞∼=H1(G) canonically. Then H1(F)∼=H1(G) as topological vector spaces by Theorem 7.9, obtaining also H1(F)∼=H1(G). /square Corollary 7.15. H2(F)andH2(F)are of finite dimension if and only if so are H2(G)andH2(G), respectively. Proof.The leaves of Gare dense since so are the leaves of F. ThusH0(G)∼=R, yieldingE·,0 2∼=H·(g)byTheorem7.1. Ontheotherhand, H1(g) =H2(g) = 0since gis compact semisimple [28]. So E1,· i=E2,· i= 0 for 2 ≤i≤ ∞by Theorem 7.1. HenceE0,2 3=E0,2 2∼=H2(F) (using Theorem 7.9), and E3,0 3=E3,0 2∼=H3(g). Therefore, since E0,2 ∞=E0,2 4= ker(d3:E0,2 3→E3,0 3), H2(G)∼=E0,2 ∞can be identified to the kernel of some continuous homomorphism of H2(F) toH3(g), and the result follows. /square Corollary 7.16. SupposeMis oriented. Let ιi:Ki→M,i= 1,2, be smooth immersions of closed oriented manifolds of complementary d imension which define homology classes of Mwith non-trivial intersection. Let Γibe the image of the composite π1(Ki)π1(ιi)− −−− →π1(M)h− −−− →G. Suppose the group generated by Γ1∪Γ2is not dense in G. If1≤k= dimK2≤2, thendimHk(F) =∞. Proof.The result follows directly applying Corollaries 7.14, 7.15, and 2.6 to G./square Corollary 7.17. SupposeMis oriented. Let ιi:Ki→M,i= 1,2, be smooth immersions of closed oriented manifolds of complementary d imension which define homology classes of Mwith non-trivial intersection. Let Γibe the image of the composite π1(Ki)π1(ιi)− −−− →π1(M)h− −−− →G. Suppose the group generated by Γ1∪Γ2is not dense in G. Ifι1is transverse to F, thendimHk(F) =∞fork= dimK2. Proof.By Corollary 7.16, we can assume k >2. LetFlH·(G) andFlH·(G),l= 0,1,2,..., be the filtrations of H·(G) andH·(G) induced by (17). We have H·(G)/F1H·(G)∼=E0,· ∞⊂E0,· 2∼=H·(F), where both isomorphisms preserve the topologies, and E0,· ∞is a closed subspace of E0,· 2. (The last isomorphism follows from Theorem 7.9.) So H·(G)/F1H·(G) can be injected into H·(F), and it is enough to prove that Hk(G)/F1Hk(G) is of infinite dimension.LEAFWISE REDUCED COHOMOLOGY 27 This is a special case of the setting of Theorem 2.2 and Corollaries 2.5 a nd 2.6. The proofs of those results yield linearly independent classes ζm∈ Hk(G). In this case, we shall prove that the ζmare also linearly independent modulo F1Hk(G). Consider the pull-back bundles ι∗ iNoverKi. The canonical maps ˜ ιi:ι∗ iN→N areimmersionstransverseto G, whichthuscanbeconsideredastransverselyregular immersionsof ( ι∗ iN,Gi) into (N,G), where Gi= ˜ι∗ iG. We can assume the ιiintersect each other transversely, and thus the ˜ ιiintersect transversely in G. Let/tildewideH·⊂H·(G) and/tildewideH·⊂ H·(G) be the subspaces given by the classes that have representatives supported in π−1 N(U) for any open subset U⊂Mcontaining ι1(K1). SetF1/tildewideH·=/tildewideH·∩F1H·(G) andF1/tildewideH·=/tildewideH·∩F1H·(G). Sinceζm∈/tildewideHkby Remark3, it is enoughto provethat the ζmarelinearly independent modulo F1/tildewideHk. Hence, according to the proof of Theorem 2.2, it is enough to prove thatι2can be chosen so that ˜ ι∗ 2/parenleftBig F1/tildewideHk/parenrightBig = 0 where ˜ ι∗ 2:H·(G)→ H·(G2). In fact we shall prove the stronger property that the choice of ι2can be made so that ˜ ι∗ 2/parenleftBig F1/tildewideHk/parenrightBig = 0 for ˜ι∗ 2:H·(G)→H·(G2). Sinceι1is transverseto F, we can choose ι2such that, for some open subset U⊂ Mcontainingι1(K1), each connected component of ι2(K2)∩Uis contained in some leaf ofF. So, for every leaf L2ofG2, the connected components of ˜ ι2(L2)∩π−1 N(U) are contained in leaves of /tildewideF, yielding ˜ι∗ 2α= 0 over˜ι−1 2π−1 N(U) for anyα∈F1Ω·(G). MoreoverUandι2can be chosen so that the connected components of ι−1 2(U) are contractible;thus˜ ι−1 2π−1 N(U)≡ι−1 2(U)×Gcanonically,wheretheslices ι−1 2(U)×{∗} are the leaves of the restriction G2,UofG2to ˜ι−1 2π−1 N(U). HenceHl(G2,U) = 0 for l>0. Finally, the abovechoices can be made so that, for some open sub setV⊂M, we haveι1(K1)∩V=∅,U∪V=M, and each connected component of ι−1 2(U∩V) is contractible. Thus, as above, Hl(G2,U∩V) = 0 forl >0, where G2,U∩Vis the restriction of G2to ˜ι−1 2π−1 N(U∩V). Therefore, by using the Mayer-Vietoris type spectral sequence (cf. [10]) ··· →Hl−1(G2,U∩V)→Hl(G2)→Hl(G2,U)⊕Hl(G2,V)→Hl(G2,U∩V)→ ··· and sincek>2, we get (23) Hk(G2)∼=Hk(G2,V) by the restriction homomorphism. Nowanyξ∈F1/tildewideHkcanbedefinedbyaleafwiseclosedform α∈Ωk(G) supported inM\Vwithα+dGβ∈F1Ωk(G) for some β∈Ωk−1(G). Then ˜ι∗ 2(α+dGβ) is supported in ˜ ι−1 2π−1 N(V), where it is the G2-leafwise derivative of ˜ ι∗ 2β. So ˜ι∗ 2ξis mapped to zeroin Hk(G2,V), and thus ˜ ι∗ 2ξ= 0 by (23), which finishes the proof. /square 8.Case of foliations on nilmanifolds Γ\Hdefined by normal subgroups of H The goal of this section is to prove Theorem 2.10. It will be done by ind uction, which needs leafwise reduced cohomology with coefficients in a vector bundle with a flatF-partial connection. Thus we shall prove a more general theorem by taking arbitrary coefficients. Forafoliation Fonamanifold Mandavectorbundle VoverM, aflatF-partial connection on Vcan be defined as a flat connection on the restriction of Vto the28 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR leaves whose local coefficients are smooth on each foliation chart of FonM. So the corresponding de Rham derivative dFwith coefficients in Vpreserves smoothness onM; i.e.dFpreservesΩ( F,V) =C∞(/logicalandtextTF∗⊗V). ThenH·(F,V) canbe defined in the same way as H·(F) by using (Ω( F,V),dF) instead of (Ω( F),dF). Consider the following particular case. Let Hbe a simply connected nilpotent Lie group, K⊂Ha normal connected subgroup, and Γ ⊂Ha discrete uniform subgroup whose projection to H/Kis dense. Then let Fbe the foliation on the nilmanifold M= Γ\Hdefined as the quotient of the foliation /tildewideFonHwhose leaves are the translates of K. In this case, Mis closed and the leaves of Fare dense. Let/tildewideVbe anH×K-vector bundle over Hfor the left action of H×KonHgiven by (h,k)h′=hh′k−1, (h,k)∈H×Kandh′∈H. We also consider the induced left actions of HandKonH. The space of H-invariant sections of /tildewideVwill be denoted by C∞/parenleftBig /tildewideV/parenrightBig H, and the subspaces of invariant sections will be denoted in a similar way for other actions. Suppose /tildewideVis endowed with an H×K-invariant flat/tildewideF-partial connection, and let Vbe the induced vector bundle on Mwith the induced flat F-partial connection. The structure of H×K-vector bundle on V canonically defines an action of konC∞/parenleftBig /tildewideV/parenrightBig H, wherekis the Lie algebra of K. Moreover the induced differential map on/logicalandtextk∗⊗C∞/parenleftBig /tildewideV/parenrightBig Hcorresponds to d/tildewideFby the canonical injection of this space in Ω( F,V). Theorem 8.1. With the above notations, H·(F,V)∼=H·/parenleftBig k,C∞/parenleftBig /tildewideV/parenrightBig H/parenrightBig . The result will follow by induction on the codimension qofF. Forq= 0 andVthe trivial line bundle, this is just a well known theorem of K. Nomizu [27]. If q= 0 andVis arbitrary, the result still follows with the obvious adaptation of the arguments in [27]. Supposeq >0 and the result is true for foliations of codimension less than q. The proof has two cases. Case1.AssumeK∩Γ = 1. ThegroupΓisnilpotentsincesois H, thusthecenterof Γ is non-trivial. Let abe a non-trivial element in the center of Γ. By the universal property of Mal’cev’s completion [23], there exists a one dimensional c onnected subgroupLof the center of Hcontaining /a\}b∇acketle{ta/a\}b∇acket∇i}htas a discrete uniform subgroup. L is isomorphic to RsinceHis simply connected. Let H1=H/L, and Γ 1= Γ//a\}b∇acketle{ta/a\}b∇acket∇i}ht. Clearly Γ 1is canonically injected in H1as a discrete uniform subgroup. We get L∩K= 1 because /a\}b∇acketle{ta/a\}b∇acket∇i}ht∩K= 1, and thus there is a canonical injection of KintoH1 as a normal subgroup, defining a foliation F1on the nilmanifold M1= Γ1\H1.F1 is a foliationofthe type considered in the statement ofthis theorem , ofcodimension q−1, but observe that the canonical injection of KintoH1may not have trivial intersection with Γ 1. The projection H//a\}b∇acketle{ta/a\}b∇acket∇i}ht →H1is canonically an S1-principal bundle (considering S1≡L//a\}b∇acketle{ta/a\}b∇acket∇i}ht), so the induced map π:M→M1is also an S1-principal bundle in a canonical way. Then Vcanonically is an S1-vector bundle so that the partial connection is invariant, and thus induces the ve ctor bundle V1=V/S1overM1with the corresponding flat F1-partial connection. The lifting ofV1toH1is/tildewideV1=/tildewideV/L, which satisfies the same properties as /tildewideVwith respect to K1instead ofK.LEAFWISE REDUCED COHOMOLOGY 29 For eachx∈M1and eachm∈Z, define Cm,x={f∈C∞(π−1(x),C) :f(yθ) =f(y)e2πmθi for ally∈π−1(x) and allθ∈S1≡R/Z}. It is easy to see that Cm=/unionsqdisplay x∈M1Cm,x is a one-dimensional C-vectorbundle over M1in a canonical way. For m∈Z, define also Ω(F,V⊗C)m={α∈Ω(F,V⊗C) :α(yθ) =α(y)e2πmθi for ally∈π−1(x) and allθ∈S1}, andsimilarlydefine C∞/parenleftBig/parenleftBig /tildewideV//a\}b∇acketle{ta/a\}b∇acket∇i}ht/parenrightBig ⊗C/parenrightBigm consideringthe S1-principalbundle H//a\}b∇acketle{ta/a\}b∇acket∇i}ht → H1. By the Fourier series expression for functions on S1, we get that Ω( F,V⊗C) is theC∞closure of /circleplusdisplay m∈ZΩ(F,V⊗C)m. It can be easily seen that there is a canonical isomorphism (24) Ω( F1,V1⊗Cm)∼=Ω(F,V⊗C)m defined byπ∗and the canonical identity C∞(Cm)≡C∞(M,C)m. SinceFis preserved by the S1-action onM,dFpreserves each Ω( F,V⊗C)mand corresponds to dF1by (24). By induction H·(F1,V1⊗Cm)∼=H·/parenleftbigg k,C∞/parenleftBig /tildewideV1⊗˜Cm/parenrightBig H1/parenrightbigg . But C∞/parenleftBig /tildewideV1⊗/tildewideCm/parenrightBig H1∼=C∞/parenleftBig/parenleftBig /tildewideV//a\}b∇acketle{ta/a\}b∇acket∇i}ht/parenrightBig ⊗C/parenrightBigm H//angbracketlefta/angbracketright canonically, which is obviously trivial if m/\e}atio\slash= 0. ButC0is the trivial complex line bundle, so H·(F,V⊗C)∼=H·(F1,V1⊗C0) ∼=H·/parenleftbigg k,C∞/parenleftBig /tildewideV1⊗C/parenrightBig H1/parenrightbigg ∼=H·/parenleftBig k,C∞/parenleftBig /tildewideV⊗C/parenrightBig H/parenrightBig . Case2.In the general case, let G=H/Kand Γ 1the projection of Γ to G. We use Mal’cev’s construction for the pair ( G,Γ1). It yields a simply connected nilpo- tent Lie group H1containing Γ 1as a discrete uniform subgroup, and a surjective homomorphism D1:H1→Gwhich is the identity on Γ 1. The kernel K1ofD1 defines a foliation Gof codimension qon the nilmanifold M1= Γ1\H1, and we have K1∩Γ1= 1. SoGis the type of foliation we have considered in Case 1. Gis the classifying foliation for foliations with transverse structure g iven by (G,Γ1). So there is a smooth map f:M→M1which is transverse to Gand so thatF=f∗G. In this particular case, fcan be constructed in the following way. By the universal property of Mal’cev’s construction, the surject ive homomorphism30 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR of Γ to Γ 1can be uniquely extended to a surjective homomorphism ˜f:H→H1, which defines a map f:M→M1. We have D1˜f=D. SoKis projected onto K1, and thus F=f∗F1. Moreover fis a locally trivial bundle with fiber the nilmanifold P/(P∩Γ), wherePis the kernel of ˜f. Fix a vector subbundle ν⊂TFwhich is complementary to the subbundle τ⊂TFof vectors that are tangent to the fibers of f. Then we get a canoni- cal isomorphism/logicalanddisplay TF∗⊗V∼=/logicalanddisplay ν∗⊗/logicalanddisplay τ∗⊗V , yielding a bigrading of Ω( F,V) given by Ωu,v(F,V) =C∞/parenleftBiggu/logicalanddisplay ν∗⊗v/logicalanddisplay τ∗⊗V/parenrightBigg . Consider the filtration of Ω( F,V) given by the differential subspaces FkΩ(F,V) =/circleplusdisplay u≥kΩu,·(F,V), which depend only on FandV; in fact they could be defined without using ν. This filtration induces a spectral sequence ( Ei,di) converging to H·(F,V), whose terms (E0,d0) and (E1,d1) can be described as follows. The derivative dFdecomposes as sum of bihomogeneous operators dF,0,1,dF,1,0anddF,2,−1, where each double subindex indicates the corresponding bidegree. These operators satisfy identities which are similar to those in (15) and (16), yielding (E0,d0)≡(Ω(F,V),dF,0,1), (E1,d1)≡(H(Ω(F,V),dF,0,1),dF,1,0∗). Letk1be the Lie algebra of K1. EachX∈k1canonically defines a vector field X1onM1which is tangent to the leaves of F1. LetXνbe the unique vector field onMwhich is a section of νand projects to X1. Forα∈Ω0,v(F) ands∈C∞(V), defineθX(α⊗s) to be the (0 ,·)-component of θXνα⊗s+α⊗∇Xνs, where∇denotes the flat F-partial connection of V. It can be easily checked that θXdF,0,1=dF,0,1θX. SoθXdefines an operator, also denoted by θX, onE0,· 1. In this way, we get a representation θofk1onE0,· 1, and a canonical isomorphism Eu,v 2∼=Hu(k1,θ). Define V1,y=H·/parenleftbig f−1(y),V|f−1(y)/parenrightbig , y∈M1, V1=/unionsqdisplay y∈M1V1,y, and let/tildewideV1be the lifting of V1toH1. It is easy to see that /tildewideV1canonically is a H1×K1-vector bundle over the H1×K1-manifoldH1with anH1×K1-invariant flat/tildewideF1-partial connection. (The fibers of /tildewideV1are of finite dimension since the fibers offare compact.) It is also easily seen that there is a canonical isomorph ism C∞(V1)∼=E0,· 1. Moreover the representation of k1onE0,· 1corresponds to the representation of k1onC∞/parenleftBig /tildewideV1/parenrightBig defined by the flat partial connection of /tildewideV1. So Eu,· 2∼=Hu/parenleftBig k1,C∞/parenleftBig /tildewideV1/parenrightBig/parenrightBig ∼=Hu(F1,V1).LEAFWISE REDUCED COHOMOLOGY 31 LetEibe the quotient of Eiover the closure 0iof its trivial subspace. Then Eu,· 2∼=Hu(F1,V1)∼=Hu/parenleftbigg k1,C∞/parenleftBig /tildewideV1/parenrightBig H1/parenrightbigg by Case 1. If the above filtration is restricted to the space of differential for ms in Ω( F,V) whose lifting to HisH-left invariant, we get a spectral sequence ( Ei,di) converging toH·/parenleftBig k1,C∞/parenleftBig /tildewideV/parenrightBig H/parenrightBig , and there is a canonical homomorphism ( Ei,di)→(Ei,di) of spectral sequences. Analogously, we have a canonical isomorp hism Eu,· 2∼=Hu(k1,C∞(V1)H1). So the composite E2→E2→ E2is an isomorphism, and thus E2∼=E2⊕02as differential complexes. Then E3∼=E3⊕H(02,d2), yielding H(02,d2)∼=03, and the above decomposition is of differential complexes. We get E4∼=E4⊕H(03,d3). Continuing with these arguments, we finally obtain Ei∼=Ei⊕0ias topological differential complexes for i≥2, and thus H·(F,V)∼=E∞∼=E∞⊕0∞. Hence H·(F,V)∼=E∞∼=H·(k,C∞(V)H) as desired. Remark 4.For general Lie foliations with dense leaves and nilpotent structura l Lie algebra,theclassifyingfoliationsareofthetypeconsideredinTheo rem2.10. Onthe one hand, if the ambient manifold is closed and the classifying map can b e chosen to be a fiber bundle, then a spectral sequence argument shows th at the leafwise reduced cohomology is of finite dimension. On the other hand, if the c lassifying map has unavoidable singularities, then they should correspond to h andles on the leavesandthe leafwisereducedcohomologyisofinfinite dimensionbyC orollary2.4. References [1] J. A. ´Alvarez L´ opez. A finiteness theorem for the spectral sequen ce of a Riemannian foliation. Illinois J. of Math. , 33:79–92, 1989. [2] J. A. ´Alvarez L´ opez. A decomposition theorem for the spectral se quence of Lie foliations. Trans. Amer. Math. Soc. , 329:173–184, 1992. [3] J. A. ´Alvarez L´ opez and G. Hector. Leafwise homologies, leafwis e cohomology, and subfo- liations. In E. Mac´ ıas-Virg´ os X. M. Masa and J. A. ´Alvarez L´ opez, editors, Analysis and Geometry in Foliated Manifolds , pages 1–12, Singapore, 1995. Proceedings of the VII Inter- national Colloquium on Differential Geometry, Santiago de C ompostela, 26–30 July, 1994, World Scientific. [4] J. A. ´Alvarez L´ opez and S. Hurder. Pure-point spectrum for folia tion geometric operators. Preprint, 1994. [5] J. A. ´Alvarez L´ opez and Y. A. Kordyukov. Long time behavior of lea fwise heat flow for Riemannian foliations. Compositio Math. , 125: 129–153, 2001. [6] P. Andrade and G. Hector. Foliated cohomology and Thurst on’s stability. Preprint, 1993. [7] P. Andrade and M. do S. Pereira. On the cohomology of one-d imensional foliated manifolds. Bol. Soc. Brasil. Mat. , 21:79–89, 1990. [8] J.L. Arraut and N. M. dos Santos. Linear foliations of Tn.Bol. Soc. Brasil. Mat. , 21:189–204, 1991. [9] R. Bott and L. W. Tu. Differential Forms in Algebraic Topology , volume 82 of Graduate Texts in Math. Springer-Verlag, New York, 1982. [10] A. El Kacimi-Alaoui. Sur la cohomologie feuillet´ ee. Compositio Math. , 49:195–215, 1983.32 J.A. ´ALVAREZ L ´OPEZ AND G. HECTOR [11] E. Fedida. Feuilletages de Lie, feuilletages du plan. I nLect. Notes Math. , volume 352, pages 183–195. Springer-Verlag, 1973. [12] C.Godbillon. Feuilletages: ´Etudes G´ eom´ etriques , volume 98 of Progress in Math. Birkh¨ auser, Boston, Basel and Stuttgart, 1991. [13] W. Greub, S. Halperin, and R. Vanstone. Connections, Curvature, and Cohomology. Vol. III. Academic Press, New York, San Francisco and London, 1975. [14] A. Haefliger. Structures feuillet´ ees et cohomologie ` a valeurs dans un faisceau de groupo¨ ıdes. Comment. Math. Helv. , 32:248–329, 1958. [15] A. Haefliger. Some remarks on foliations with minimal le aves.J. Differential Geom. , 15:269– 384, 1980. [16] A. Haefliger. Leaf closures in Riemannian foliations. I nA Fˆ ete on Topology , pages 3–32, New York, 1988. Academic Press. [17] G. Hector and U. Hirsch. Introduction to the Geometry of Foliations, Part B , volume E3 of Aspects of Mathematics . Friedr. Vieweg and Sohn, Braunschweig, 1983. [18] J. L. Heitsch. A cohomology for foliated manifolds. Comment. Math. Helv. , 50:197–218, 1975. [19] S. Hurder. Spectral theory of foliation geometric oper ators. Preprint, 1992. [20] A. El Kacimi-Alaoui and A. Tihami. Cohomologie bigradu ´ ee de certains feuilletages. Bull. Soc. Math. Belge , s´ erie B-38, 1986. [21] W. Klingenberg. Riemannian Geometry , volume 1 of De Gruyter Studies in Mathematics . De Gruyter, Berlin, 1982. [22] S. Kobayashi. On conjugate and cut loci. In Studies in Global Geometry and Analysis , pages 96–122. Math. Assoc. of Amer., 1967. [23] A.I. Mal’cev. On a class of homogeneous spaces. Transl. Amer. Math. Soc. , 39:276–307, 1951. [24] P. Molino. G´ eom´ etrie globale des feuilletages riema nniens.Nederl. Akad. Wetensch. Indag. Math., 44:45–76, 1982. [25] P. Molino. Riemannian Foliations , volume 73 of Progress in Math. Birkh¨ auser Boston Inc., Boston, MA, 1988. [26] P. Molino and M. Pierrot. Th´ eor´ emes des slice et holon omie des feuilletages Riemanniens. Annales Inst. Fourier, Grenoble , 37:207–223, 1987. [27] K. Nomizu. On the cohomology of compact homogeneous spa ces of nilpotent lie groups. Ann. of Math. , 59:531–538, 1954. [28] W. A. Poor. Differential Geometric Structures . McGraw-Hill, New York, 1982. [29] B. L. Reinhart. Foliated manifolds with bundle-like me trics.Ann. of Math. , 69:119–132, 1959. [30] C. Roger. M´ ethodes Homotopiques et Cohomologiques en Th´ eorie de Fe uilletages . Universit´ e de Paris XI, Paris, 1976. [31] J. Stallings. Group Theory and Three Dimensional Manifolds . Yale University Press, Yale, 1971. [32] R. Thom. Quelques propri´ et´ es globales des vari´ et´ e s diff´ erentiables. Comment. Math. Helv. , 28:17–86, 1954. Departamento de Xeometr ´ıa e Topolox ´ıa, Facultade de Matem ´aticas, Universidade de Santiago de Compostela, 15782 Santiago de Compostela, Sp ain E-mail address :jesus.alvarez@usc.es Institut Girard Desargues, UPRESA 5028, 43, boulevard du 11 No vembre 1918, Uni- versit´e Claude Bernard-Lyon I, 69622 Villeurbanne Cedex, France E-mail address :hector@geometrie.univ-lyon1.fr | 2013-11-14 | Geometric conditions are given so that the leafwise reduced cohomology is of
infinite dimension, specially for foliations with dense leaves on closed
manifolds. The main new definition involved is the intersection number of
subfoliations with "appropriate coefficients". The leafwise reduced cohomology
is also described for homogeneous foliations with dense leaves on closed
nilmanifolds. | The dimension of the leafwise reduced cohomology | 1311.3518v1 |
arXiv:2210.16931v1 [math.AP] 30 Oct 2022Intrinsic polynomial squeezing for Balakrishnan-Taylor beam models E. H. Gomes Tavares, M. A. Jorge Silva, V. Narciso and A. Vicen te Abstract We explore the energy decay properties related to a model in e xtensible beams with the so-called energy damping . We investigate the influence of the non- loncal damping coefficient in the stability of the model. We pr ove, for the first time, that the corresponding energy functional is squeezed by pol ynomial-like functions involving the power of the damping coefficient, which arises i ntrinsically from the Balakrishnan-Taylor beam models. As a consequence, it is sh own that such models with nonlocal energy damping are never exponentially stabl e in its essence. 1 Introduction In 1989 Balakrishnan and Taylor [3] derived some prototypes of vibrating exten- sible beams with the so-called energy damping . Accordingly, the following one dimensional beam equation is proposed /u1D715/u1D461/u1D461/u1D462−2/u1D701√ /u1D706/u1D715/u1D465/u1D465/u1D462+/u1D706/u1D715/u1D465/u1D465/u1D465/u1D465/u1D462−/u1D6FC/bracketleftbigg/uni222B.dsp/u1D43F −/u1D43F/parenleftbig/u1D706|/u1D715/u1D465/u1D465/u1D462|2+ |/u1D715/u1D461/u1D462|2/parenrightbig/u1D451/u1D465/bracketrightbigg/u1D45E /u1D715/u1D465/u1D465/u1D461/u1D462=0,(1) where/u1D462=/u1D462(/u1D465,/u1D461)represents the transversal deflection of a beam with length 2 /u1D43F >0 in the rest position, /u1D6FC > 0 is a damping coefficient, /u1D701is a constant appearing in Krylov-Bogoliubov’s approximation, /u1D706 > 0 is related to mode frequency and spectral density of external forces, and /u1D45E=2(/u1D45B+/u1D6FD) +1 with/u1D45B∈Nand 0≤/u1D6FD<1 2. E. H. Gomes Tavares State University of Londrina, 86057-970, Londrina, PR, Bra zil, e-mail:eduardogomes7107@gmail.com M. A. Jorge Silva State University of Londrina, 86057-970, Londrina, PR, Bra zil. e-mail:marcioajs@uel.br V. Narciso State University of Mato Grosso do Sul, 79804-970, Dourados , MS, Brazil. e-mail:vnarciso@uems.br A. Vicente Western Paraná State University, 85819-110, Cascavel, PR, Brazil. e-mail:andre.vicente@unioeste.br 12 E. H. Gomes Tavares, M. A. Jorge Silva, V. Narciso and A. Vice nte We still refer to [3, Sect. 4] for several other beam equation s taking into account nonlocal energy damping coefficients, as well as [2, 4, 6, 7, 12 , 17, 18] for associated models. A normalized /u1D45B-dimensional equation corresponding to (1) can be seen as follows /u1D715/u1D461/u1D461/u1D462−/u1D705Δ/u1D462+Δ2/u1D462−/u1D6FC/bracketleftbigg/uni222B.dsp Ω/parenleftBig |Δ/u1D462|2+ |/u1D715/u1D461/u1D462|2/parenrightBig /u1D451/u1D465/bracketrightbigg/u1D45E Δ/u1D715/u1D461/u1D462=0, (2) where we denote /u1D706=1 and/u1D705=2/u1D701;Ωmay represent an open bounded of R/u1D45B; and the symbols ΔandΔ2stand for the usual Laplacian and Bi-harmonic operators, respectively. Additionally, in order to see the problem wit hin the frictional context of dampers, we rely on materials whose viscosity can be essen tially seen as friction between moving solids. In this way, besides reflecting on a mo re challenging model (at least) from the stability point of view, one may metaphys ically supersede the viscous damping in (2) by a nonlocal frictional one so that we cast the model /u1D715/u1D461/u1D461/u1D462−/u1D705Δ/u1D462+Δ2/u1D462+/u1D6FC/bracketleftbigg/uni222B.dsp Ω/parenleftBig |Δ/u1D462|2+ |/u1D715/u1D461/u1D462|2/parenrightBig /u1D451/u1D465/bracketrightbigg/u1D45E /u1D715/u1D461/u1D462=0. (3) The main goal of this paper is to explore the influence of the no nloncal damping coefficient in the stability of problem (3). Unlike the existi ng literature on extensible beams with full viscous or frictional damping, we are going t o see for the first time that the feature of the energy damping coefficient E/u1D45E(/u1D461):=E/u1D45E(/u1D462,/u1D462/u1D461)(/u1D461)=/bracketleftbigg/uni222B.dsp Ω/parenleftBig |Δ/u1D462(/u1D461)|2+ |/u1D715/u1D461/u1D462(/u1D461)|2/parenrightBig /u1D451/u1D465/bracketrightbigg/u1D45E , /u1D45E > 0, (4) not only prevents exponential decay, but also gives us a poly nomial range in terms of/u1D45Ewhose energy is squeezed and goes to zero polynomially when t ime goes to infinity. More precisely, by noting that the corresponding e nergy functional is given by /u1D438/u1D705(/u1D461):=/u1D438/u1D705(/u1D462,/u1D462/u1D461)(/u1D461)=/uni222B.dsp Ω/parenleftBig |Δ/u1D462(/u1D461)|2+ |/u1D715/u1D461/u1D462(/u1D461)|2+/u1D705|∇/u1D462(/u1D461)|2/parenrightBig /u1D451/u1D465, /u1D705≥0,(5) then it belongs to an area of variation between upper and lowe r polynomial limits as follows /u1D4500/u1D461−1 /u1D45E/lessorsimilar/u1D438/u1D705(/u1D461)/lessorsimilar/u1D4360/u1D461−1 /u1D45E, /u1D461→ +∞, (6) for some constants 0 < /u1D450 0≤/u1D4360depending on the initial energy /u1D438/u1D705(0), /u1D705≥0. Indeed, such a claim corresponds to an intrinsic polynomial range of (uniform) stability and will follow as a consequence of a more general r esult that is rigorous stated in Theorem 2. See also Corollary 1. In particular, we c an conclude that (3) is not exponentially stable when dealing with weak initial dat a, that is, with solution in the standard energy space. See Corollary 2. In conclusion, Theorem 2 truly reveals the stability of the a ssociated energy /u1D438/u1D705(/u1D461), which leads us to the concrete conclusions provided by Coro llaries 1-2, being pioneering results on the subject. Due to technicalities in the well-posedness process,Intrinsic polynomial squeezing for Balakrishnan-Taylor b eam models 3 we shall work with /u1D45E≥1/2. In Section 2 we prepare all notations and initial results. Then, all precise details on the stability results shall be g iven in Section 3. 1.1 Previous literature, comparisons and highlights In what follows, we are going to highlight that our approach a nd results are different or else provide generalized results, besides keeping more p hysical consistency in working exactly with (4) instead of modified versions of it. I ndeed, there are at least three mathematical ways of attacking the energy damping coe fficient (4) along the equation (3) (or (2)), namely: 1. Keeping the potential energy in (4), but neglecting the ki netic one; 2. Keeping the kinetic energy in (4), but neglecting the pote ntial one; 3. Keeping both potential and kinetic energies, but conside ring them under the action of a strictly (or not) positive function /u1D440(·)as a non-degenerate (or possibility degenerate) damping coefficient. In the first case, equation (3) becomes to /u1D715/u1D461/u1D461/u1D462−/u1D705Δ/u1D462+Δ2/u1D462+/u1D6FC/bracketleftbigg/uni222B.dsp Ω|Δ/u1D462|2/u1D451/u1D465/bracketrightbigg/u1D45E /u1D715/u1D461/u1D462=0 inΩ× (0,∞). (7) This is, for sure, the most challenging case once the damping coefficient becomes now to a real degenerate coefficient. In [5, Theorem 3.1], work ing on a bounded domainΩwith clamped boundary condition, it is proved the following with/u1D45E=1 in (7): for every/u1D445 >0, there exist constants /u1D436/u1D445=/u1D436(/u1D445)>0and/u1D6FE/u1D445=/u1D6FE(/u1D445)>0 depending on /u1D445such that /u1D438/u1D705(/u1D461) ≤/u1D436/u1D445/u1D438/u1D705(0)/u1D452−/u1D6FE/u1D445/u1D461, /u1D461 > 0, (8) only holds for every regular solution /u1D462of(3)with initial data (/u1D4620,/u1D4621)satisfying /ba∇dbl(/u1D4620,/u1D4621)/ba∇dbl(/u1D43B4(Ω)∩/u1D43B2 0(Ω))×/u1D43B2 0(Ω)≤/u1D445. (9) We stress that (8) only represents a local stability result since it holds on every ball with radius /u1D445 > 0 in the strong topology (/u1D43B4(Ω) ∩/u1D43B2 0(Ω)) ×/u1D43B2 0(Ω),but they are not independent of the initial data. Moreover, as ob served by the authors in [5], the drawback of (8)-(9) is that it could not be proved i n the weak topology /u1D43B2 0(Ω) ×/u1D43F2(Ω), even taking initial data uniformly bounded in /u1D43B2 0(Ω) ×/u1D43F2(Ω). Although we recognized that our results for (3) can not be fai rly compared to such a result, we do can conclude by means of the upper and lowe r polynomial bounds (6) that the estimate (8) will never be reached for wea k initial data given in /u1D43B2 0(Ω) ×/u1D43F2(Ω). Therefore, our results act as complementary conclusions t o [5] by clarifying such drawback raised therein, and yet giving a di fferent point of view of stability by means of (6) and its consequences concerning pr oblem (3).4 E. H. Gomes Tavares, M. A. Jorge Silva, V. Narciso and A. Vice nte In the second case, equation (3) falls into /u1D715/u1D461/u1D461/u1D462−/u1D705Δ/u1D462+Δ2/u1D462+/u1D6FC/bracketleftbigg/uni222B.dsp Ω|/u1D715/u1D461/u1D462|2/u1D451/u1D465/bracketrightbigg/u1D45E /u1D715/u1D461/u1D462=0 inΩ× (0,∞). (10) Unlike the first case, here we have an easier setting because t he kinetic damping coefficient provides a kind of monotonous (polynomial) dampi ng whose computa- tions to achieve (6) remain unchanged (and with less calcula tions). This means that all results highlighted previously still hold for this part icular case. In addition, they clarify what is precisely the stability result related to pr oblems addressed in [19, 20], which in turn represent particular models of abstract dampi ng given by [1, Section 8]. In other words, in terms of stability, our methodology pr ovides a way to show the existence of absorbing sets with polynomial rate (and no t faster than polynomial rate depending on /u1D45E) when dealing with generalized problems relate to (10), sub ject that is not addressed in [19, 20]. Finally, in the third case let us see equations (2)-(3) as fol lows /u1D715/u1D461/u1D461/u1D462−/u1D705Δ/u1D462+Δ2/u1D462+/u1D440/parenleftbigg/uni222B.dsp Ω/parenleftBig |Δ/u1D462|2+ |/u1D715/u1D461/u1D462|2/parenrightBig /u1D451/u1D465/parenrightbigg /u1D434/u1D715/u1D461/u1D462=0 inΩ× (0,∞),(11) where operator /u1D434represents the Laplacian operator /u1D434=−Δor else the identity one /u1D434=/u1D43C. Thus, here we clearly have two subcases, namely, when /u1D440(·) ≥ 0 is a non- degenerate or possibly degenerate function. For instance, when/u1D440(/u1D460)=/u1D6FC/u1D460/u1D45E, /u1D460≥0, and/u1D434=−Δ, then we go back to problem (2). For this (degenerate) nonloc al strong damping situation with /u1D45E≥1, it is considered in [11, Theorem 3.1] an upper polynomial stability for the corresponding energy, which a lso involves a standard nonlinear source term. Nonetheless, we call the attention t o the following prediction result provided in [11, Theorem 4.1] for (2) addressed on a bo unded domain Ωwith clamped boundary condition and /u1D45E≥1:By taking finite initial energy 0</u1D438/u1D705(0)< ∞, then/u1D438/u1D705(/u1D461)given in (5)satisfies /u1D438/u1D705(/u1D461) ≤3/u1D438/u1D705(0)/u1D452−/u1D6FF∫/u1D461 0/ba∇dbl/u1D462(/u1D460)/ba∇dbl2/u1D45E/u1D451/u1D460, /u1D461 > 0, (12) where/u1D6FF=/u1D6FF(1 /u1D438/u1D705(0))>0is a constant proportional to 1//u1D438/u1D705(0). Although the estimate (12) provides a new result with an exponential face , it does not mean any kind of stability result. Indeed, it is only a peculiar estimate indicating that prevents exponential decay patterns as rem arked in [11, Section 4]. In addition, it is worth pointing out that our computations t o reach the stability result for problem (3) can be easily adjusted to (2), even for /u1D45E≥1/2 thanks to a inequality provided in [1, Lemma 2.2]. Therefore, through the polynomi al range (6) we provide here a much more accurate stability result than the estimate expressed by (12), by concluding indeed that both problems (2) and (3) are never ex ponentially stable in the topology of the energy space. On the other hand, in the non-degenerate case /u1D440(/u1D460)>0, /u1D460≥0, but still taking /u1D434=−Δ, a generalized version of (11) has been recently approached by [16] in a context of strong attractors , that is, the existence of attractors in the topology ofIntrinsic polynomial squeezing for Balakrishnan-Taylor b eam models 5 more regular space than the weak phase space. In this occasio n, the/u1D4361-regularity for/u1D440 > 0 brings out the non-degeneracy of the damping coefficient, wh ich in turn allowed them to reach interesting results on well-posednes s, regularity and long-time behavior of solutions over more regular spaces. Such assump tion of positiveness for the damping coefficient has been also addressed by other au thors for related problems, see e.g. [8, 9, 10]. From our point of view, in spite of representing a nice case, the latter does not portray the current situation of th is paper so that we do not provide more detailed comparisons with such a non-degenera te problems, but we refer to [5, 8, 9, 10, 11, 16] for a nice survey on this kind of no n-degenerate damping coefficients. Additionally, we note that the suitable case of non-degenerate damping coefficient/u1D440(/u1D460)>0, /u1D460≥0, and/u1D434=/u1D43Cin (11) has not been considered in the literature so far and shall be concerned in another work by th e authors in the future. At light of the above statements, one sees e.g. when /u1D440(/u1D460)=/u1D6FC/u1D460/u1D45E, /u1D460≥0,and /u1D434=/u1D43C, then problem (11) falls into (3), being a problem not yet add ressed in the literature that brings out a new branch of studies for such a n onlocal (possibly degenerate) damped problems, and also justifies all new stab ility results previously specified. 2 The problem and well-posedness Let us consider again the beam model with energy damping /u1D715/u1D461/u1D461/u1D462+Δ2/u1D462−/u1D705Δ/u1D462+/u1D6FC/bracketleftbigg/uni222B.dsp Ω/parenleftBig |/u1D715/u1D461/u1D462|2+ |Δ/u1D462|2/parenrightBig /u1D451/u1D466/bracketrightbigg/u1D45E /u1D715/u1D461/u1D462=0 inΩ× (0,∞),(13) with clamped boundary condition /u1D462=/u1D715/u1D462 /u1D715/u1D708=0 on/u1D715Ω× [0,∞), (14) and initial data /u1D462(/u1D465,0)=/u1D4620(/u1D465), /u1D715/u1D461/u1D462(/u1D465,0)=/u1D4621(/u1D465), /u1D465∈Ω. (15) To address problem (13)-(15), we introduce the Hilbert phas e space (still called energy space ) H:=/u1D43B2 0(Ω) ×/u1D43F2(Ω), equipped with the inner product/angbracketleftbig /u1D4671,/u1D4672/angbracketrightbig H:=/angbracketleftbig Δ/u1D4621,Δ/u1D4622/angbracketrightbig +/angbracketleftbig /u1D463.alt1,/u1D463.alt2/angbracketrightbig for/u1D467/u1D456=(/u1D462/u1D456,/u1D463.alt/u1D456) ∈ H, /u1D456=1,2,and norm /ba∇dbl/u1D467/ba∇dblH=/parenleftbig/ba∇dblΔ/u1D462/ba∇dbl2+ /ba∇dbl/u1D463.alt/ba∇dbl2/parenrightbig1/2,for/u1D467=(/u1D462,/u1D463.alt) ∈ H,where /an}b∇acke⊔le{⊔/u1D462,/u1D463.alt/an}b∇acke⊔∇i}h⊔:=/uni222B.dsp Ω/u1D462/u1D463.alt/u1D451/u1D465 ,/ba∇dbl/u1D462/ba∇dbl2:=/an}b∇acke⊔le{⊔/u1D462,/u1D462/an}b∇acke⊔∇i}h⊔and/ba∇dbl/u1D467/ba∇dbl2 H:=/an}b∇acke⊔le{⊔/u1D467,/u1D467/an}b∇acke⊔∇i}h⊔H. In order to stablish the well-posedness of (13)-(15), we defi ne the vector-valued function/u1D467(/u1D461):=(/u1D462(/u1D461),/u1D463.alt(/u1D461)),/u1D461≥0,with/u1D463.alt=/u1D715/u1D461/u1D462. Then we can rewrite system (13)-(15) as the following first order abstract problem6 E. H. Gomes Tavares, M. A. Jorge Silva, V. Narciso and A. Vice nte /braceleftBigg/u1D715/u1D461/u1D467=A/u1D467+M(/u1D467), /u1D461 > 0, /u1D467(0)=(/u1D4620,/u1D4621):=/u1D4670,(16) whereA:D(A) ⊂ H → H is the linear operator given by A/u1D467=(/u1D463.alt,−Δ2/u1D462),D(A) :=/u1D43B4(Ω) ∩/u1D43B2 0(Ω), (17) andM:H → H is the nonlinear operator M(/u1D467)=(0,/u1D705Δ/u1D462−/u1D6FC/ba∇dbl/u1D467/ba∇dbl2/u1D45E H/u1D463.alt), /u1D467=(/u1D462,/u1D463.alt) ∈ H. (18) Therefore, the existence and uniqueness of solution to the s ystem (13)-(15) relies on the study of problem (16). Accordingly, we have the follow ing well-posedness result. Theorem 1. Let/u1D705,/u1D6FC≥0and/u1D45E≥1 2be given constants. If /u1D4670∈ H, then (16)has a unique mild solution /u1D467in the class/u1D467∈/u1D436([0,∞),H). In addition, if /u1D4670∈ D(A) , then/u1D467is a regular solution lying in the class /u1D467∈/u1D436([0,∞),D(A)) ∩/u1D4361([0,∞),H). Proof. To show the local version of the first statement, it is enough t o prove that A given in (17) is the infinitesimal generator of a /u1D4360-semigroup of contractions /u1D452A/u1D461 (which is very standard) and Mset in (18) is locally Lipschitz on Hwhich will be done next. Indeed, let /u1D45F >0 and/u1D4671,/u1D4672∈ Hsuch that max {/ba∇dbl/u1D4671/ba∇dblH,/ba∇dbl/u1D4672/ba∇dblH} ≤/u1D45F. We note that /bardblex/bardblex/bardblex/ba∇dbl/u1D4671/ba∇dbl2/u1D45E H/u1D463.alt1− /ba∇dbl/u1D4672/ba∇dbl2/u1D45E H/u1D463.alt2/bardblex/bardblex/bardblex≤/bracketleftBig /ba∇dbl/u1D4671/ba∇dbl2/u1D45E H+ /ba∇dbl/u1D4672/ba∇dbl2/u1D45E H/bracketrightBig /ba∇dbl/u1D463.alt1−/u1D463.alt2/ba∇dbl+/barex/barex/barex/ba∇dbl/u1D4671/ba∇dbl2/u1D45E H− /ba∇dbl/u1D4672/ba∇dbl2/u1D45E H/barex/barex/barex/ba∇dbl/u1D463.alt1+/u1D463.alt2/ba∇dbl. (19) The first term on the right side of (19) can be estimated by /bracketleftBig /ba∇dbl/u1D4671/ba∇dbl2/u1D45E H+ /ba∇dbl/u1D4672/ba∇dbl2/u1D45E H/bracketrightBig /ba∇dbl/u1D463.alt1−/u1D463.alt2/ba∇dbl ≤2/u1D45F2/u1D45E/ba∇dbl/u1D4671−/u1D4672/ba∇dblH. Now, from a suitable inequality provided in [1] /one.supwe estimate the second term as follows /barex/barex/barex/ba∇dbl/u1D4671/ba∇dbl2/u1D45E H− /ba∇dbl/u1D4672/ba∇dbl2/u1D45E H/barex/barex/barex/ba∇dbl/u1D463.alt1+/u1D463.alt2/ba∇dbl ≤4/u1D45E/u1D45F2/u1D45E/ba∇dbl/u1D4671−/u1D4672/ba∇dblH. Plugging the two last estimates in (19), we obtain /bardblex/bardblex/bardblex/ba∇dbl/u1D4671/ba∇dbl2/u1D45E H/u1D463.alt1− /ba∇dbl/u1D4672/ba∇dbl2/u1D45E H/u1D463.alt2/bardblex/bardblex/bardblex H≤2(2/u1D45E+1)/u1D45F2/u1D45E/ba∇dbl/u1D4671−/u1D4672/ba∇dblH. Thus, /one.supSee [1, Lemma 2.2]: Let/u1D44Bbe a normed space with norm /ba∇dbl · /ba∇dbl/u1D44B. Then, for any /u1D460≥1we have /barex/barex/ba∇dbl/u1D462/ba∇dbl/u1D460 /u1D44B− /ba∇dbl/u1D463.alt/ba∇dbl/u1D460 /u1D44B/barex/barex≤/u1D460max{/ba∇dbl/u1D462/ba∇dbl/u1D44B,/ba∇dbl/u1D463.alt/ba∇dbl/u1D44B}/u1D460−1/ba∇dbl/u1D462−/u1D463.alt/ba∇dbl/u1D44B,∀/u1D462,/u1D463.alt∈/u1D44B. (20)Intrinsic polynomial squeezing for Balakrishnan-Taylor b eam models 7 /ba∇dblM(/u1D4671) −M(/u1D4672)/ba∇dblH≤/parenleftBig /u1D705+2(2/u1D45E+1)/u1D6FC/u1D45F2/u1D45E/parenrightBig /ba∇dbl/u1D4671−/u1D4672/ba∇dblH, andMis locally Lipschitz in H. Hence, according to Pazy [15, Chapter 6], if /u1D4670∈ H(/u1D4670∈/u1D437(A)), there exists a time/u1D461max∈ (0,+∞]such that (16) has a unique mild (regular) solution /u1D467∈/u1D436([0,/u1D461max),H) (/u1D467∈/u1D436([0,/u1D461max),/u1D437(A)) ∩/u1D4361([0,/u1D461max),H)). Moreover, such time /u1D461maxsatisfies either the conditions /u1D461max=+∞or else/u1D461max<+∞ with lim /u1D461→/u1D461−max/ba∇dbl/u1D467(/u1D461)/ba∇dblH=+∞. (21) In order to show that /u1D461max=+∞, we consider /u1D4670∈/u1D437(A)and the corresponding regular solution /u1D467of (16). Taking the inner product in Hof (16) with /u1D467, we obtain 1 2/u1D451 /u1D451/u1D461/bracketleftbig /ba∇dbl/u1D467(/u1D461)/ba∇dbl2 H+/u1D705/ba∇dbl∇/u1D462(/u1D461)/ba∇dbl2/bracketrightbig +/u1D6FC/ba∇dbl/u1D467(/u1D461)/ba∇dbl2/u1D45E H/ba∇dbl/u1D715/u1D461/u1D462(/u1D461)/ba∇dbl2=0/u1D461∈ [0,/u1D461max).(22) Integrating (22) over (0,/u1D461), /u1D461∈ [0,/u1D461max), we get /ba∇dbl/u1D467(/u1D461)/ba∇dblH≤ (1+/u1D450′/u1D705)1/2/ba∇dbl/u1D4670/ba∇dblH, /u1D461∈ [0,/u1D461max). Here, the constant /u1D450′>0 comes from the embedding /u1D43B2 0(Ω)↩→/u1D43B1 0(Ω). The last estimate contradicts (21). Hence, /u1D461/u1D45A/u1D44E/u1D465=+∞. Using a limit process, one can conclude the same result for mild solutions. The proof of Theorem 1 is then complete. 3 Lower-upper polynomial energy’s bounds By means of the notations introduced in Section 2, we recall t hat the energy functional corresponding to problem (13)-(15) can be expressed by /u1D438/u1D705(/u1D461)=1 2/bracketleftbig /ba∇dbl(/u1D462(/u1D461),/u1D715/u1D461/u1D462(/u1D461))/ba∇dbl2 H+/u1D705/ba∇dbl∇/u1D462(/u1D461)/ba∇dbl2/bracketrightbig , /u1D461≥0. (23) Our main stability result reveals that /u1D438/u1D705(/u1D461)is squeezed by decreasing polynomial functions as follows. Theorem 2. Under the assumptions of Theorem 1, there exists an increasing function J:R+→R+such that the energy /u1D438/u1D705(/u1D461)satisfies /bracketleftbig 2/u1D45E+1/u1D6FC/u1D45E/u1D461+/bracketleftbig /u1D438/u1D705(0)/bracketrightbig−/u1D45E/bracketrightbig−1//u1D45E≤/u1D438/u1D705(/u1D461) ≤/bracketleftbigg/u1D45E J (/u1D438/u1D705(0))(/u1D461−1)++/bracketleftbig /u1D438/u1D705(0)/bracketrightbig−/u1D45E/bracketrightbigg−1//u1D45E ,(24) for all/u1D461 >0, where we use the standard notation /u1D460+:=(/u1D460+ |/u1D460|)/2. Proof. Taking the scalar product in /u1D43F2(Ω)of (13) with /u1D715/u1D461/u1D462, we obtain8 E. H. Gomes Tavares, M. A. Jorge Silva, V. Narciso and A. Vice nte /u1D451 /u1D451/u1D461/u1D438/u1D705(/u1D461)=−/u1D6FC||(/u1D462(/u1D461),/u1D715/u1D461/u1D462(/u1D461))||2/u1D45E H/ba∇dbl/u1D715/u1D461/u1D462(/u1D461)/ba∇dbl2, /u1D461 > 0. (25) Let us prove the lower and upper estimates in (24) in the seque l. Lower bound. We first note that ||(/u1D462(/u1D461),/u1D715/u1D461/u1D462(/u1D461))||2/u1D45E H/ba∇dbl/u1D715/u1D461/u1D462(/u1D461)/ba∇dbl2≤2/u1D45E+1[/u1D438/u1D705(/u1D461)]/u1D45E+1, and replacing it in (25), we get /u1D451 /u1D451/u1D461/u1D438/u1D705(/u1D461) ≥ − 2/u1D45E+1/u1D6FC[/u1D438/u1D705(/u1D461)]/u1D45E+1, /u1D461 > 0. (26) Thus, integrating (26) and proceeding a straightforward co mputation, we reach the first inequality in (24). Upper bound. Now, we are going to prove the second inequality of (24). To do so, we provide some proper estimates and then apply a Nakao’s res ult (cf. [13, 14]). We start by noting that ||(/u1D462(/u1D461),/u1D715/u1D461/u1D462(/u1D461)||2/u1D45E H/ba∇dbl/u1D715/u1D461/u1D462(/u1D461)/ba∇dbl2≥ /ba∇dbl/u1D715/u1D461/u1D462(/u1D461)/ba∇dbl2(/u1D45E+1), (27) and replacing (27) in (25), we get /u1D451 /u1D451/u1D461/u1D438/u1D705(/u1D461) +/u1D6FC/ba∇dbl/u1D715/u1D461/u1D462(/u1D461)/ba∇dbl2(/u1D45E+1)≤0, /u1D461 > 0, (28) which implies that /u1D438/u1D705(/u1D461)is non-increasing with /u1D438/u1D705(/u1D461) ≤/u1D438/u1D705(0)for every/u1D461 >0. Also, integrating (28) from /u1D461to/u1D461+1, we obtain /u1D6FC/uni222B.dsp/u1D461+1 /u1D461/ba∇dbl/u1D715/u1D461/u1D462(/u1D460)/ba∇dbl2(/u1D45E+1)/u1D451/u1D460≤/u1D438/u1D705(/u1D461) −/u1D438/u1D705(/u1D461+1):=[/u1D437(/u1D461)]2. (29) Using Hölder’s inequality with/u1D45E /u1D45E+1+1 /u1D45E+1=1 and (29), we infer /uni222B.dsp/u1D461+1 /u1D461/ba∇dbl/u1D715/u1D461/u1D462(/u1D460)/ba∇dbl2/u1D451/u1D460≤1 /u1D6FC1 /u1D45E+1[/u1D437(/u1D461)]2 /u1D45E+1. (30) From the Mean Value Theorem for integrals, there exist /u1D4611∈ [/u1D461,/u1D461+1 4]and/u1D4612∈ [/u1D461+3 4,/u1D461+1]such that /ba∇dbl/u1D715/u1D461/u1D462(/u1D461/u1D456)/ba∇dbl2≤4/uni222B.dsp/u1D461+1 /u1D461/ba∇dbl/u1D715/u1D461/u1D462(/u1D460)/ba∇dbl2/u1D451/u1D460≤4 /u1D6FC1 /u1D45E+1[/u1D437(/u1D461)]2 /u1D45E+1, /u1D456=1,2. (31) On the other hand, taking the scalar product in /u1D43F2(Ω)of (13) with /u1D462and inte- grating the result over [/u1D4611,/u1D4612], we haveIntrinsic polynomial squeezing for Balakrishnan-Taylor b eam models 9 /uni222B.dsp/u1D4612 /u1D4611/u1D438/u1D705(/u1D460)/u1D451/u1D460=/uni222B.dsp/u1D4612 /u1D4611/ba∇dbl/u1D715/u1D461/u1D462(/u1D460)/ba∇dbl2/u1D451/u1D460+1 2[(/u1D715/u1D461/u1D462(/u1D4611),/u1D462(/u1D4611)) − (/u1D715/u1D461/u1D462(/u1D4612),/u1D462(/u1D4612))] −/u1D6FC 2/uni222B.dsp/u1D4612 /u1D4611||(/u1D462(/u1D460),/u1D715/u1D461/u1D462(/u1D460))||2/u1D45E H(/u1D715/u1D461/u1D462(/u1D460),/u1D462(/u1D460))/u1D451/u1D460. (32) Let us estimate the terms in the right side of (32). Firstly, w e note that through Hölder’s inequality, (31) and Young’s inequality, we obtai n |(/u1D715/u1D461/u1D462(/u1D4611),/u1D462(/u1D4611)) − (/u1D715/u1D461/u1D462(/u1D4612),/u1D462(/u1D4612))| ≤/u1D4512/summationdisplay.1 /u1D456=1/ba∇dbl/u1D715/u1D461/u1D462(/u1D461/u1D456)/ba∇dbl/ba∇dblΔ/u1D462(/u1D461/u1D456)/ba∇dbl ≤8/u1D451 /u1D6FC1 2(/u1D45E+1)[/u1D437(/u1D461)]1 /u1D45E+1sup /u1D4611≤/u1D460≤/u1D4612[/u1D438/u1D705(/u1D460)]1/2 ≤128/u1D4512 /u1D6FC1 /u1D45E+1[/u1D437(/u1D461)]2 /u1D45E+1+1 8sup /u1D4611≤/u1D460≤/u1D4612/u1D438/u1D705(/u1D460), where the constant /u1D451 >0 comes from the embedding /u1D43B2 0(Ω)↩→/u1D43F2(Ω). Addition- ally, using that /u1D438/u1D705(/u1D461) ≤/u1D438/u1D705(0), we have /ba∇dbl(/u1D462(/u1D461),/u1D715/u1D461/u1D462(/u1D461))/ba∇dbl2/u1D45E H≤2/u1D45E[/u1D438/u1D705(/u1D461)]/u1D45E≤2/u1D45E[/u1D438/u1D705(0)]/u1D45E. From this and (30) we also get /barex/barex/barex/barex/uni222B.dsp/u1D4612 /u1D4611||(/u1D462(/u1D460),/u1D715/u1D461/u1D462(/u1D460))||2/u1D45E H(/u1D715/u1D461/u1D462(/u1D460),/u1D462(/u1D460))/u1D451/u1D460/barex/barex/barex/barex≤22/u1D45E+3/u1D4512[/u1D438/u1D705(0)]2/u1D45E /u1D6FC−/u1D45E /u1D45E+1[/u1D437(/u1D461)]2 /u1D45E+1 +1 8/u1D6FCsup /u1D4611≤/u1D460≤/u1D4612/u1D438/u1D705(/u1D460). Regarding again (30) and replacing the above estimates in (3 2), we obtain /uni222B.dsp/u1D4612 /u1D4611/u1D438/u1D705(/u1D460)/u1D451/u1D460≤ K (/u1D438/u1D705(0)) [/u1D437(/u1D461)]2 /u1D45E+1+1 8sup /u1D4611≤/u1D460≤/u1D4612/u1D438/u1D705(/u1D460), (33) where we set the function Kas K(/u1D460):=/bracketleftbigg64/u1D4512+1 /u1D6FC1 /u1D45E+1+2(/u1D45E+1)/u1D4512/u1D6FC2/u1D45E+1 /u1D45E+1/u1D4602/u1D45E/bracketrightbigg >0. Using once more the Mean Value Theorem for integrals and the f act that/u1D438/u1D705(/u1D461)is non-increasing, there exists /u1D701∈ [/u1D4611,/u1D4612]such that /uni222B.dsp/u1D4612 /u1D4611/u1D438/u1D705(/u1D460)/u1D451/u1D460=/u1D438/u1D705(/u1D701)(/u1D4612−/u1D4611) ≥1 2/u1D438/u1D705(/u1D461+1), and then10 E. H. Gomes Tavares, M. A. Jorge Silva, V. Narciso and A. Vic ente sup /u1D461≤/u1D460≤/u1D461+1/u1D438/u1D705(/u1D460)=/u1D438/u1D705(/u1D461)=/u1D438/u1D705(/u1D461+1) + [/u1D437(/u1D461)]2≤2/uni222B.dsp/u1D4612 /u1D4611/u1D438/u1D705(/u1D460)/u1D451/u1D460+ [/u1D437(/u1D461)]2. Thus, from this and (33), we arrive at sup /u1D461≤/u1D460≤/u1D461+1/u1D438/u1D705(/u1D460) ≤ [/u1D437(/u1D461)]2+2/uni222B.dsp/u1D4612 /u1D4611/u1D438/u1D705(/u1D460)/u1D451/u1D460 ≤ [/u1D437(/u1D461)]2+2K (/u1D438/u1D705(0)) [/u1D437(/u1D461)]2 /u1D45E+1+1 4sup /u1D461≤/u1D460≤/u1D461+1/u1D438/u1D705(/u1D460), and since 0<2 /u1D45E+1≤2, we obtain sup /u1D461≤/u1D460≤/u1D461+1/u1D438/u1D705(/u1D460) ≤4 3[/u1D437(/u1D461)]2 /u1D45E+1/bracketleftBig [/u1D437(/u1D461)]2/u1D45E /u1D45E+1+2K (/u1D438/u1D705(0))/bracketrightBig . (34) Observing that [/u1D437(/u1D461)]2/u1D45E /u1D45E+1≤ [/u1D438/u1D705(/u1D461) +/u1D438/u1D705(/u1D461+1)]/u1D45E /u1D45E+1≤2/u1D45E /u1D45E+1[/u1D438/u1D705(0)]/u1D45E /u1D45E+1,and de- noting by J(/u1D460):=/parenleftbigg4 3/parenrightbigg/u1D45E+1/bracketleftBig (2/u1D460)/u1D45E /u1D45E+1+2K(/u1D460)/bracketrightBig/u1D45E+1 >0, (35) and also recalling the definition of [/u1D437(/u1D461)]2in (29), we obtain from (34) that sup /u1D461≤/u1D460≤/u1D461+1[/u1D438/u1D705(/u1D460)]/u1D45E+1≤ J (/u1D438/u1D705(0)) [/u1D438/u1D705(/u1D461) −/u1D438/u1D705(/u1D461+1)]. Hence, applying e.g. Lemma 2.1 of [14] with /u1D438/u1D705=/u1D719,J (/u1D438/u1D705(0))=/u1D4360,and/u1D43E=0, we conclude /u1D438/u1D705(/u1D461) ≤/bracketleftbigg /u1D45E J(/u1D438/u1D705(0))(/u1D461−1)++1/bracketleftbig /u1D438/u1D705(0)/bracketrightbig/u1D45E/bracketrightbigg−1//u1D45E ,which ends the proof of the second inequality in (24). The proof of Theorem 2 is therefore complete. Remark 1. It is worth point out that we always have /bracketleftbig 22/u1D45E+1/u1D6FC/u1D45E/u1D461+/bracketleftbig /u1D438/u1D705(0)/bracketrightbig−/u1D45E/bracketrightbig−1//u1D45E≤/bracketleftbigg/u1D45E J (/u1D438/u1D705(0))(/u1D461−1)++/bracketleftbig /u1D438/u1D705(0)/bracketrightbig−/u1D45E/bracketrightbigg−1//u1D45E ,(36) so that it makes sense to express /u1D438/u1D705(/u1D461)between the inequalities in (24). Indeed, from the definition Jin (35) one easily sees that J (/u1D438/u1D705(0)) ≥1 22/u1D45E+1/u1D6FC,from where one concludes (36) promptly. Corollary 1. (Polynomial Range of Decay). Under the assumptions of Theorem 2, the energy functional /u1D438/u1D705(/u1D461)defined in (23)decays squeezed as follows /u1D4500/u1D461−1 /u1D45E/lessorsimilar/u1D438/u1D705(/u1D461)/lessorsimilar/u1D4360/u1D461−1 /u1D45Eas/u1D461→ +∞, (37) for some constants 0</u1D4500≤/u1D4360depending on the initial energy /u1D438/u1D705(0).Intrinsic polynomial squeezing for Balakrishnan-Taylor b eam models 11 In other words, /u1D438/u1D705(/u1D461)decays polynomially at rate /u1D461−1//u1D45E(/u1D45E≥1/2) as/u1D461→ +∞ . ⊓ ⊔ Corollary 2. (Non-Exponential Stability). Under the assumptions of Theorem 2, the energy/u1D438/u1D705(/u1D461)set in (23)never decays exponentially as /u1D452−/u1D44E/u1D461(/u1D44E >0) as/u1D461→ +∞ . ⊓ ⊔ References 1. F. Aloui, I. Ben Hassen, A. Haraux, Compactness of traject ories to some nonlinear second order evolution equations and applications, J. Math. Pures Appl. 100 (2013), no. 3, 295-326. 2. A. V. Balakrishnan, A theory of nonlinear damping in flexib le structures. Stabilization of flexible structures, p. 1-12, 1988. 3. A. V. Balakrishnan and L. W. Taylor, Distributed paramete r nonlinear damping models for flight structures, in: Proceedings Daming 89, Flight Dynami cs Lab and Air Force Wright Aeronautical Labs, WPAFB, 1989. 4. R. W. Bass and D. Zes, Spillover, Nonlinearity, and flexibl e structures, in The Fourth NASA Workshop on Computational Control of Flexible Aerospace Sy stems, NASA Conference Publication 10065 ed. L.W.Taylor (1991) 1-14. 5. M. M. Cavalcanti, V. N. Domingos Cavalcanti, M. A. Jorge Si lva, V. Narciso, Stability for extensible beams with a single degenerate nonlocal damping of Balakrishnan-Taylor type, J. of Differential Equations, 290 (2021), 197-222. 6. E. H. Dowell, Aeroelasticity of plates and shells, Gronin ger, NL, Noordhoff Int. Publishing Co. (1975). 7. T. J. Hughes and J. E. Marsden, Mathematical foundation of elasticity, Englewood C. Prentice- Hall, 1983. 8. M. A. Jorge Silva and V. Narciso, Long-time behavior for a p late equation with nonlocal weak damping, Differential Integral Equations 27 (2014), no. 9-1 0, 931-948. 9. M. A. Jorge Silva and V. Narciso, Attractors and their prop erties for a class of nonlocal extensible beams, Discrete Contin. Dyn. Syst. 35 (2015) no. 3, 985-1008. 10. M. A. Jorge Silva and V. Narciso, Long-time dynamics for a class of extensible beams with nonlocal nonlinear damping, Evol. Equ. Control Theory 6 (20 17), no. 3, 437-470. 11. M. A. Jorge Silva, V. Narciso and A. Vicente, On a beam mode l related to flight structures with nonlocal energy damping, Discrete and Continuous Dyna mical Systems Series B. 24 (2019), 3281-3298. 12. C. Mu, J. Ma, On a system of nonlinear wave equations with B alakrishnan-Taylor damping, Z. Angew. Math. Phys. 65 (2014) 91-113. 13. M. Nakao, Convergence of solutions of the wave equation w ith a nonlinear dissipative term to the steady state, Mem. Fac. Sci. Kyushu Univ. Ser. A, 30 (19 76), 257-265. 14. M. Nakao, A difference inequality and its application to n onlinear evolution equations, J. Math. Soc. Japan 30 4 (1978) 747-762. 15. A. Pazy, Semigroups of linear operators and application s to partial differential equations, vol. 44, Springer-Verland,1983. 16. Y. Sun, Z. Yang, Strong attractors and their robustness f or an extensible beam model with energy damping, Discrete and Continuous Dynamic al Systems, 2021. (doi: 10.3934/dcdsb.2021175) 17. Y. You, Inertial manifolds and stabilization of nonline ar beam equations with Balakrishnan- Taylor damping, Abstr. Appl. Anal. 1(1) (1996) 83-102. 18. W. Zhang, Nonlinear Damping Model: Response to Random Ex citation in: 5th Annual NASA Spacecraft Control Laboratory Experiment (SCOLE) Worksho p (1988) 27-38.12 E. H. Gomes Tavares, M. A. Jorge Silva, V. Narciso and A. Vic ente 19. C. Zhao, C. Zhao, C. Zhong, The global attractor for a clas s of extensible beams with nonlocal weak damping, Discrete Contin. Dyn. Syst. - B, 25 (2020), 935 -955. 20. C. Zhao, S. Ma, C. Zhong, Long-time behavior for a class of extensible beams with nonlocal weak damping and critical nonlinearity Journal of Mathemat ical Physics 61 (2020), p. 032701. | 2022-10-30 | We explore the energy decay properties related to a model in extensible beams
with the so-called energy damping. We investigate the influence of the
nonloncal damping coefficient in the stability of the model. We prove, for the
first time, that the corresponding energy functional is squeezed by
polynomial-like functions involving the power of the damping coefficient, which
arises intrinsically from the Balakrishnan-Taylor beam models. As a
consequence, it is shown that such models with nonlocal energy damping are
never exponentially stable in its essence. | Intrinsic polynomial squeezing for Balakrishnan-Taylor beam models | 2210.16931v1 |
Transport properties of spin superfluids—comparing easy-plane ferro- and antiferromagnets Martin Evers and Ulrich Nowak Fachbereich Physik, Universität Konstanz, 78457 Konstanz, Germany (Dated: December 2, 2019) We present a study on spin-superfluid transport based on an atomistic, classical spin model. Easy- plane ferro- as well as antiferromagnets are considered, which allows for a direct comparison of these two material classes based on the same model assumptions. We find a spin-superfluid transport which is robust against variations of the boundary conditions, thermal fluctuations, and dissipation modeled via Gilbert damping. Though the spin accumulations is smaller for antiferromagnets the range of the spin-superfluid transport turns out to be identical for ferro- and antiferromagnets. Fi- nally, we calculate and explore the role of the driving frequency and especially the critical frequency, where phase slips occur and the spin accumulation breaks down. I. INTRODUCTION Spin transport in magnetic insulators [1, 2] has been intensively studied beacause of the fundamental interest in the various physical phenomena that occur in these materials and because of their potential for future appli- cations. Magnetic insulators do not exhibit Joule heat- ing [3] as no electron transport is involved and many of these are oxides with exceptionally low magnetic damp- ing [4], which hopefully allows for energy efficient trans- port properties. It has even been shown that the realiza- tion of logic elements is possible [5], such that devices are compatible and integratable with CMOS technology [6]. Studies on transport in this material class focuses mostly on transport of magnons [7], i.e. quanta of spin waves— the elementary excitations of the magnetic ground state. As magnons are quasi particles, their number is not con- served and each magnon mode shows an exponential de- cay upon transport through the system on a length scale ξcalled magnon propagation length [8–13]. This is even true at zero temperature and in a clean system without any disorder due to the coupling of the magnons to elec- tronic and phononic degrees of freedom, a fact which is described phenomenologically via Gilbert damping in the equation of motion as will be explained below. In contrast to this damped magnonic transport, a pro- posal for spin transport was made that carries the name spin superfluidity. The original idea is in fact quite old [14, 15] and rests on a similarity of the magnetic or- der parameter—either the magnetization of a ferromag- net or the Néel vector of an antiferromagnet—compared to the order parameter of superfluidity—the macroscopic wave function—as it occurs for He-4 below the lambda transition. For instance, in easy-plane ferromagnets the magnetizationfeaturesaspontaneouslybrokenrotational symmetry in the easy plane ( SO(2)symmetry) that is equivalent to the spontaneously broken gauge invariance of the macroscopic wave function ( U(1)symmetry). This symmetry leads in both cases to currents that are sta- ble against small deviations—the supercurrents. [16] One striking difference of spin-superfluid transport to spin- wave transport is its distance dependence: for spin su-perfluidity it is expected to be non-exponential, pushing the limit of the range of magnonic transport. The first experimental realizations of a spin superfluid was achieved in a system of nuclear spins of He-3 atoms [17]—a model system which is not in a solid state. Only recently the physics of spin superfluidity has drawn again attention for the case of solid magnets [18–23], including a proposed dissipationless transport in metallic magnets [18]. However, König et al. neglected spin-orbit inter- action in their model for the electrons, which is one of the reasons for Gilbert damping in magnets [24]. But ev- ery known material exhibits spin-orbit interaction—since spinandangularmomentumofanatomareneverexactly zero—and therefore also magnetic damping, even if it is small. Consequently, spinsuperfluidsdoalwaysshowdis- sipation in contrast to their conventional counterparts. Recent theoretical work has focused on insulators rather than metals, usually based of phenomenological modelsincluding theLandau-Lifshitz-Gilbertequationof motion for both ferro- and antiferromagnets. [16, 19, 20] The experimental detection of spin superfluidity in solid- state magnets has been reported for magnon condensates [25], where the origin of the spin-superfluid order param- eter is different to the cases described above, and also in antiferromagnetic solids [23]. However, the interpre- tation of the experimental findings is still controversially discussed [16, 26–28]. In the following, we will investigate and compare spin superfluidityinferro-andantiferromagneticmodels. The geometry of our model resembles that of an experimen- tal non-local spin-transport investigation as sketched in fig. 1. In the corresponding experiments [29] at one side (here on the left) a spin current is injected into the mag- net viathe spin-Halleffect causedby an electricalcurrent through an attached heavy-metal stripe. The resulting spin current is detected using the inverse spin-Hall ef- fect at another position (here the right-hand side). In our model we avoid the details of the excitation mech- anism and model the effect of the injected spin current by an appropriate boundary condition that triggers the dynamics of the spin systems that we investigate. This is done from the perspective of an atomistic, classicalarXiv:1911.12786v1 [cond-mat.mes-hall] 28 Nov 20192 Figure 1. Basic concept of non-local spin transport as in an experimental setup: heavy metal stripes are attached to the magnet to inject a spin current via the spin-Hall effect (here on the left hand side). The spin current in a certain distance (here at the right end) is detected via inverse spin-Hall effect. spin model, which has some advantages: the approach is not restricted to small deviations from the ground state, finite temperatures can be investigated and our calcu- lations are not limited to the steady state only. Fur- thermore, we are able to compare ferro- and antiferro- magnetic systems. Their behavior turns out to be very similar, except for the resulting spin accumulation that is muchlowerforthelatter. However,fromanexperimental point of view antiferromagnets are much more promising, since these are not prone to a breakdown of spin super- fluidity as a consequence of dipolar interactions, which is hard to avoid in ferromagnets. [22] II. ATOMISTIC SPIN MODEL We consider the following classical, atomistic spin model of Heisenberg type [30], comprising Nnormal- ized magnetic moments Sl=µl/µSon regular lattice sitesrl. We assume a simple cubic lattice with lattice constanta. The Hamiltonian for these moments, in the following called “spins”, is given by H=−J 2/summationdisplay /angbracketleftn,m/angbracketrightSn·Sm−dz/summationdisplay n(Sn z)2,(1) taking into account Heisenberg exchange interaction of nearest neighbors quantified by the exchange constant J, where each spin has Nnbnearest neighbors. Further- more, a uniaxial anisotropy with respect to the zdirec- tion with anisotropy constant dzis included. In this work we consider the easy-plane case dz<0, where the mag- nets ground state readsgSl=±(cos(gϕ),sin(gϕ),0)with some arbitrary, but uniform anglegϕ∈[0,2π](SO(2) symmetry) and an alternating sign ±in case of antifer- romagnetic order ( J <0). The time evolution of the spins Slis governed by the stochastic Landau-Lifshitz-Gilbert (LLG) equationof motion [31–33] dSl dt=−γ µS(1 +α2)/bracketleftbig Sl×/parenleftbig Hl+αSl×Hl/parenrightbig/bracketrightbig (2) Hl=−∂H ∂Sl+ξl /angbracketleftbig ξl β(t)/angbracketrightbig = 0,/angbracketleftBig ξl β(t)ξl/prime η(t/prime)/angbracketrightBig =δll/primeδβηδ(t−t/prime)2µSαkBT γ describing the motion of a spin in its effective field Hl, whereγisthegyromagneticratio, αtheGilbertdamping constant,kBthe Boltzmann constant and Tthe absolute temperature. The properties of the thermal noise ξlare chosen such that the dissipation-fluctuation theorem is satisfied [34]. The material parameters define our sys- tem of units,|J|for the energy, tJ:=µS/γ|J|for the time,afor the distance. Numerically the LLG equation is solved either by the classical Runge-Kutta method in case of zero temperature, or at finite temperature using stochastic Heun’s method. At zero temperature the dis- sipated power per spin due to Gilbert damping follows directly from the time evolution of the spins Sl(t)[35]: Pdiss=1 NdH dt=1 N/summationdisplay n∂H ∂Sn/bracehtipupleft/bracehtipdownright/bracehtipdownleft/bracehtipupright eff.field·∂Sn ∂t/bracehtipupleft/bracehtipdownright/bracehtipdownleft/bracehtipupright LLG.(3) We study a magnetic wire extended along xdirection. The system size for our numerical simulations is given byN=Nx×Ny×Nzspins along x-,y- andzdirec- tion, where Nx/greatermuchNy,Nz. For transverse directions we use periodic boundary conditions if not noted otherwise. Boundary spins at x=Nxa(the right-hand side) are denotedSl/vextendsingle/vextendsingle rightand at this side an open boundary con- dition is applied, Sl/vextendsingle/vextendsingle right= 0. At the opposite side, at x= 0, we use a time-dependent boundary condition, Sl/vextendsingle/vextendsingle left=±(cos(ω0t),sin(ω0t),0), (4) in form of an in-plane precession with frequency ω0that injects a spin current from this side. The alternating sign (±) is used only for antiferromagnetic systems, according to the sublattices with antiparallel spin orientation. The use of this boundary condition creates an ex- citation with well-defined frequency ω0. Alternatively, we also assumed an externally given spin accumulation µ=µezat the left-hand side that causes additional torques on the spins and drives them out of equilibrium, which directly maps an experimental implementation us- ing a spin-Hall-generated spin accumulation to the model utilized here. This method has been used for instance in [22]. In appendix B we calculate how this spin accumula- tion maps to the excitation frequency ω0and we further- more confirmed numerically that both mechanisms lead to the same response for ferro- and antiferromagnets. Although an atomistic picture—comprising discrete degrees of freedom—is studied numerically, the micro- magnetic approximation is of particular value for analyt- ical considerations of ferromagnets. This approximation3 assumes that spatial variations of magnetic structures are small compared to the atomic distance a. In this case differences can be approximated as derivatives and the spins form a continuous field S(r,t). It is handy to use cylindrical coordinates S=/parenleftBig/radicalbig 1−S2zcosϕ,/radicalbig 1−S2zsinϕ, Sz/parenrightBig ,where definitions Sz(rl) :=Sl zandϕ(rl) :=ϕllink the atomistic picture to the micromagnetics. Note that for a spin superfluid Szis considered as the spin-superfluid density and ϕits phase. The use of the micromagnetic approximationforferromagnetsallowstoreformulatethe LLGequationintermsofdifferentialequationsfor Szand ϕthat read µS γ˙ϕ=Ja2/bracketleftBigg 1 1−S2z∆Sz+Sz|∇Sz|2 (1−S2z)2+Sz|∇ϕ|2/bracketrightBigg + 2dzSz−αµS γ˙Sz 1−S2z(5) µS γ˙Sz=−Ja2/bracketleftbig/parenleftbig 1−S2 z/parenrightbig ∆ϕ−2Sz∇Sz·∇ϕ/bracketrightbig +α/parenleftbig 1−S2 z/parenrightbigµS γ˙ϕ. (6) These two equations are strictly equivalent to the LLG equation eq. (2) for zero temperature with the only as- sumption of the micromagnetic approximation. If one expands these equations in lowest order in ∇ϕ,∆ϕ,∇Sz, and∆Szfor an easy-plane magnet, which implies espe- cially assuming|Sz|/lessmuch1, but keeping|∇ϕ|2, one ends up with µS γ˙ϕ=Ja2∆Sz+Ja2Sz|∇ϕ|2+ 2dzSz−αµS γ˙Sz(7) µS γ˙Sz=−Ja2∆ϕ+αµS γ˙ϕ. (8) Importantly, keeping the |∇ϕ|2term is actually required if the damping takes relatively high values, a fact which we checked numerically. Furthermore, these equations are very similar to others already reported in [19, 21], but not exactly equivalent. Ref. [19] uses more approxima- tions, especially neglecting the |∇ϕ|2-term, and ref. [21] considers a different starting point, namely a quantum theory at low temperatures, where this term has a dif- ferentSz-dependence. Because of this difference, the re- sult from [21] does not exactly match our numerical re- sults of the atomistic spin model, nor does it match the classical micromagnetic theory. Hence, we use eqs. (7) and (8) that do describe the atomistic spin simulations well. However, eqs. (7) and (8) can be solved in steady state for a special case: a ferromagnet that is of length L alongxdirection and exhibits translational invariance in y- andzdirection as carried out in appendix A. Steady state means a coherent precession of all spins with a fre- quency ˙ϕ=ω0and a stationary profile Sz(x). This so- lution of eqs. (7) and (8) reads: sϕ(x,t) =α 2µSω0 γJ(x−L)2 a2+ω0t+ϕ0(9) sSz(x) =sSz(L) 1 +µ2 Sω2 0 2γ2Jdzα2/parenleftbigx−L a/parenrightbig2, (10) with a spin accumulation at the right end of the sys- tem (atx=Nxa=:L) ofsSz(L) =µSω0/2γdz, avalue which is independent of L—one of the striking fea- tures of spin superfluidity. Another feature is the mono- tone increase of ϕwhich implies the formation of an in- plane spin spiral with winding number Nw, which reads 2πNw=/integraltext dϕ=ϕ(L)−ϕ(0). Note furthermore, that an open boundary condition at the right end is an assump- tion that leads to solutions eqs. (9) and (10), correspond- ing to a Neumann condition ∇ϕ/vextendsingle/vextendsingle right= 0, which must be justified as a realistic choice. For the numerical study of eq. (2) we assume an open boundary at the right end. Equation (10) assumes the same and results in a finite Szatx= 0, which contra- dicts the numerical driving boundary at this side, eq. (4), that forces Sz(x=0) = 0. Furthermore, in an experiment an open boundary at the right end might not be feasi- ble because of outflowing spin currents, for example into an attached heavy metal. Thus, the real behavior at the boundaries for sure deviates from the ideal solution eq. (10) and raises the question how strong that devia- tion is and in how far the boundary conditions influence the overall bulk behavior of the spin transport. This is examined numerically from the full model eq. (2) by varying the boundary conditions on the left and right. One example of the variations we tested is an absorbing boundary condition on the right, modeling an outflow- ing spin current by an enhanced damping. As result we observe the profile Sl zto show only little change in that case compared to an open boundary and also that in all cases the numerical profiles well follow eq. (10) (see in the following fig. 2 a) as example). Other variations of the boundary condition which we tested have also hardly any impact on the magnets overall response. III. EASY-PLANE FERROMAGNET In a first step of the numerical investigation, we con- sider a collinear ferromagnet as most simple case, with parameters J > 0for the ferromagnetic state and dz= −0.01Jas in-plane anisotropy. Let us describe the phe-4 0 1000 2000 3000 4000 5000012345610-3 010203040506070 0 0.5 1 1.5 2 2.5 10-300.010.020.03 5 6 10-40.0240.026 Figure 2. Spin superfluidity in a 1D ferromagnet at T= 0in the steady state: a)depicts the spin accumulation Szand the in-plane angle ϕforω0tJ=−2×10−4; numerical data (blue and red symbols) follow perfectly the theoretical curve eqs. (9) and (10) (black, dashed lines), except for the vicinity of the left boundary. This is an artifact of the boundary condition, eq. (4), used for the numerics. b)shows the spin accumulation at the right end of the system SN zversus driving frequency ω0; for small driving frequencies up to a critical value ωcritthe numerical data follow the analytical curvesSz(L); for larger frequencies the spin accumulation breaks down, deviating form the theoretical curve, due to phase slips and spin wave excitations. nomenology of the spin superfluid in a 1D system of size Nx×Ny×Nz= 5000×1×1at temperature T= 0. This model is equivalent to a 3D system with transla- tional invariance in y- andzdirection. Furthermore, we setα= 0.05andω0tJ=−2×10−4. Starting from a uniform ferromagnet as initial condi- tion, the boundary spin starts to rotate and due to ex- change the next spin will follow this rotation and ac- cordingly drive its neighbor and so on. But since a spin cannot immediately follow the dynamics of its neighbor, there is a certain phase difference Dϕbetween the spins, i.e., the neighbor to the right is lagging behind. In the micromagnetic approximation this effect is described by a phase gradient ∇ϕ≈Dϕ/a. The rotation of the spins speeds up, until it reaches the final precession frequency, given by the driving frequency ω0. At the same time theout-of-planecomponent Sz—thespinaccumulation— increases until it reaches a steady state profile. The time scale of this transient phase for reaching a steady state can be quantified: ˙ϕ(t)andSz(t)follow a limited expo- nential growth on a characteristic time τt≈5×105tJ for the parameters used here. τtscales positively with system size Nxand damping α. Eventually, the numerical time evolution reaches a steady state as shown in fig. 2 a). This steady state verifies the analytical solution eqs. (9) and (10) in bulk with a deviation only at the left boundary as anticipated and described above. Note that the finite spin accumu- lationSzas a consequence of this type of dynamics has importantfeatures: itisalong-rangespintransportsince it decays non-exponential and it would allow to measure spin transport by means of the inverse spin-Hall effect. Furthermore, it could also be addressed, for instance, by magneto-optical measurements—if sensitive to the out- of-plane magnetization for a geometry as studied here. For a further investigation, we vary the frequency ω0 and find two different regimes, one for sufficiently smallω0, where the system is able to follow the excitation without disturbance, and one for large ω0where the sys- tems response deviates from the theoretical expectation. Thesetworegimes, whichwewillcalllinearandnonlinear regime in the following, are sharply separated by a crit- ical frequency ωcrit. The existence of these two regimes can be seen from the data depicted in fig. 2 b). Here, as a measure, we consider the spin accumulation of the last spinSN zat the right end of the system. Below ωcritwe find just the analytical valuesSz(L), see eq. (10), which scales linearly with ω0. Atωcritthis behavior breaks down and the spin accumulation SN zdecreases with in- creasing pumping frequency. This breakdown can be un- derstood in terms of the phase gradient ∇ϕwhich scales linearly with the driving frequency ω0, see eq. (9). How- ever, one can expect a maximum phase gradient ∇ϕfor a spin-superfluid state given by the Landau criterion [36]: if the phase gradient exceeds locally a critical value, it is energetically favorable for the spins at this position to rotate out of the x-yplane and return to the plane by unwinding the spiral. Hence, the winding number Nw decreases by one—an effect which is called a phase slip. The Landau criterion for the stability of a spin superfluid with respect to phase slips reads [36] |∇ϕ|</radicalbigg −2dz Ja2. (11) Note that this relation is not exact as a uniform Szis assumed for its derivation. Nevertheless do we observe these phase slips numerically. In the linear regime the winding number is constant in the steady state, whereas in the nonlinear regime it relaxes by one at a regular rate Γpsas shown in fig. 3 at the example of ω0tJ= −6.5×10−4. Theω0dependence of the phase-slip rate Γpsis de- picted in fig. 4. Each phase slip is accompanied by the excitation of a broad spin-wave spectrum on top of the5 4 4.005 4.01 4.015 4.02 10748.54949.550 00.511.510-7 Figure 3. Winding number Nwand dissipated power Pdiss in the steady state of a driven ferromagnet for ω0tJ= −6.5×10−4, well in the nonlinear regime. At a rate of Γps phase slips relax the winding number of the in-plane spiral. For each such event the dissipated power spikes. spin superfluid. These spin waves are visible as oscil- lations ofSzaround the spin-superfluid magnitude and, hence, there is strictly speaking no steady state any more as the phase slips and the spin-wave excitation are def- initely time dependent. In particular, for systems with low Gilbert damping this dynamics leads to determinis- ticchaos, thoughthe spin-superfluidbackground remains visible. These findings have some severe implications as there is a maximum spin accumulation, which is achieved right at the edge between the linear and the nonlinear regime. Furthermore, driving the system in the nonlin- ear regime means also to waste energy to the phase slips and the excitation of incoherent spin waves. For the parameters here the critical frequency takes the valueωcrittJ≈−5.15×10−4, which is determined from fig. 2 b). We also tested different parameters, vary- ingαandL(data not shown), and find that ωcritscales negatively with αandL. Our numerical result can be compared to the analytical prediction above, eq. (11). From eq. (9) follows that the maximum phase gradient is given by∇sϕ(0) =αµSω0L/γJa2, which, inserted into eq. (11), implies |ωcrit|=γJa αµSL/radicalbigg −2dz J. (12) For our parameter set this takes value 4×10−41/tJ, which is slightly lower compared to the numerical value above. This discrepancy is probably due to the fact that eq. (11) ignores the spatial dependence of Sz. Further- more, a test for very low damping α= 10−4showed that eq. (11) is even more inaccurate in that case. Another important quantity is the dissipated power given by eq. (3), which takes negative sign as it lowers the total energy. Figure 4 depicts its dependence on ω0. Belowthecriticalfrequency, inthesteadystate, itistime independent as the dynamics is completely stationary. In this regime it scales quadratically with the excitation frequency,Pdiss∝ω2 0, a result which has already been 10-410-310-210-1010-910-810-7 10-610-410-2Figure 4. Time-averaged dissipated power /angbracketleftPdiss/angbracketrightand phase- slip rate Γpsin the steady state versus driving frequency ω0comparing ferro- (FM) and antiferromagnets (AFM). The perpendicular dash-dotted lines mark ωcritfor the FM and the AFM, where the latter takes on the higher value. For ω0< ω critthe dissipated power scales as /angbracketleftPdiss/angbracketright∝ω2 0and is identical for FMs and AFMs. Above ωcritthe increase slows down and the curve flattens for very high ω0. In this regime, /angbracketleftPdiss/angbracketrightis higher for AFMs as compared to FMs. For the phase-slip rate we find Γps= 0for|ω0|<|ωcrit|and similar, increasing values for the FM and the AFM above ωcrit. The deviation between FM and AFM near ωcritis due to the dif- ferent critical frequency, i.e. the data almost coincide when plotted versus ω0−ωcrit. reported before [37]. This behavior changes above ωcrit. Thetimeevolutionofthedissipationinthisregimeshows that the phase slips notably contribute to the dissipated power, i.e. for each phase slip Pdissspikes as shown in fig. 3. Because of this time-dependence of Pdiss, we have to consider an average over time for the evaluation of the dissipated power. Still, the dissipated power increases furtherwith ω0butlessthanlinearandthecurvenotably flattens. So far, our results were obtained from zero- temperature simulations. In the following we address the robustness of spin-superfluid transport at finite tem- perature. For this, we consider a finite cross section Ny×Nz>1andNx= 2000 and vary the tempera- ture. An average over Navrealizations of thermal noise is carried out and, furthermore, data are averaged over thecrosssectioninordertoreducethenoise. Thespecific choice of parameters in provided is table I. Figure 5 presents the numerical results for the exam- ple ofkBT/J = 10−2forSzandϕ. The spin-superfluid transport remains in tact but, in particular, the spin ac- cumulation Szshows strong thermal fluctuations despite the averages taken over the cross section and the Navre- alizations. However, on average the spin accumulation clearly deviates from its equilibrium value, which is zero. To quantify the influence of the temperature we calculate the spatial average over the xdirection/angbracketleftSz/angbracketrightxand com- parethistothezero-temperaturevalue, givenbyeq.(10). The results are included in Table I. Furthermore, the in- plane angle/angbracketleftϕ/angbracketrightNavshows only little fluctuations and its6 00.020.04 0 500 1000 1500 200001020 Figure 5. Spin superfluidity in a ferromagnet at finite tem- peraturekBT/J = 10−2and forω0tJ=−2×10−4: shown is the spin accumulation Szand the in-plane angle ϕ. Blue lines represent the numerical data, black dash-dotted lines the an- alytical solution at zero-temperature. The spin accumulation is subjected to strong thermal fluctuations but still has a fi- niteaveragevalue /angbracketleftSz/angbracketrightx=/summationtext nSn z/Nx, depictedasreddashed line. Its value is only slightly lower than the zero-temperature value. Thermal fluctuations are much less pronounced for the in-plane angle. Table I. Averaged spin accumulation of a ferromagnet driven withω0tJ=−2×10−4for different temperatures. The cor- responding zero-temperature value is /angbracketleftSz/angbracketright= 0.01, from which no significant deviation is observed. kBT/JNx×NyNav/angbracketleftSz/angbracketrightx 10−44×4 38 0.010 10−24×4 15 0.009 0.05 8×8 5 0.010 0.10 8×8 4 0.011 0.20 14×145 0.012 spatial profile shows hardly any deviation from the zero- temperature behavior, given by eq. (9). Overall, we find no significant difference to the zero temperature case. We also checked whether phase slips due to thermal ac- tivation can be observed, but from the available data we could not observe a single one with the conclusion that ΓpstJ<4×10−5. Hence, spin superfluidity is very robust against thermal fluctuations, even though these fluctuations are a problem in our simulations in terms of the signal-to-noise ratio. IV. EASY-PLANE ANTIFERROMAGNETS For antiferromagnets, the magnetic unit cells comprise twoatoms—denotedAandBinthefollowing—thatform two sublattices. We write all properties using this label- ing so thatASlandBSlare spins of the corresponding sublattices. In the ground state both sublattices have opposite orientation,ASl=−BSl. The field equations, eqs. (5) and (6), do not hold as these require a small in-plane angle difference between two neighboring spins Dϕ, which is obviously not true in this case. However,it is reasonable to define phase differences and gradients within each sublattice, i.e.ADϕas phase difference be- tween a spin of sublattice A and its next-nearest neigh- bor, which is the nearest neighbor within sublattice A. Accordingly,BDϕdefines the phase difference of sublat- ticeB. Assuming sufficiently weak excitation, spatial variationswithineachsublatticearesmallsuchthatami- cromagnetic approximation inside the sublattices reads ∇A,Bϕ≈A,BDϕ/2a. Interestingly, numerical results re- veal that the antiferromagnetic system in bulk fulfills field equation (8), applied separately to each sublattice. The other eq. (7) is not valid, as has been reported before [20] for a phenomenological model for antiferromagnets. Consequently, the antiferromagnet is expected to exhibit the same in-plane angleA,Bϕ(up to phase difference of π betweensublattices)asaferromagnetwithcorresponding parameters, but not the same spin accumulationA,BSz. Before we discuss the numerical results in detail, let us first introduce two differences to the ferromagnet that are essential for understanding the following results: the roleofexchangeand(interlinkedwiththis)thetransverse geometry. Just as in a ferromagnet, a spin-superfluid dy- namics imposes a finite spin accumulationA,BSzwhich, remarkably, carries the same sign for both sublattices leading to a small out-of-plane magnetization. But this is of course antagonized by the antiferromagnetic exchange that favors antiparallel orientation of all components be- tween sublattices. Consequently, the exchange interac- tions must lower the spin accumulation Sztremendously as compared to the ferromagnet (compare fig. 6 a) and fig. 2 a)). This also implies that the behavior of Szis determined by the number of nearest neighbors Nnbof a spin as more neighbors imply stronger exchange cou- pling. Consequently, a 1D spin chain is less prone to this exchange reduction than a 3D system. We checked this numerically by comparing 1D, 2D and 3D models and, indeed, the spin accumulation of the spin superfluid Sz scales linearly with Nnb. There is another implication: at a boundary the num- berofneighborsislocallyreduced—andthereforetheim- portance of the exchange—, resulting in deviations of the sublattice componentsA,BSz, see fig. 6 a) for a 1D setup (the effect is less pronounced in 3D). This 1D setup owns only boundaries along the xdirection and the question whether for finite cross section Ny×Nz>1these devi- ations aty- andzboundaries significantly influence the bulk behavior has also been tested numerically. Fortu- nately, deviations at transversal boundaries quickly fall off with distance to the boundary over a few lattice con- stants. The bulk then behaves qualitatively and quan- titatively just as a 1D system, except for the reduced spin accumulation due to the number of neighbors as discussed above. The study of 1D systems is preferable to keep computational costs feasible. We turn now to the presentation of the numerical data for a 1D system. The model parameters are the same as given above for the ferromagnet, except for the exchange constant which is now negative. Similarly to the ferro-7 050100024610-5 10002000300040004900 5000 0 0.5 1 1.5 2 2.5 10-300.511.510-4 5.566.5 10-41.31.41.510-4 Figure 6. Spin superfluidity in antiferromagnetic spin chains: a)the spin accumulation in the stead state resolved for the two sublattices A and B. In the bulk both take the same value, leading to a finite total spin accumulation, which is two orders of magnitude lower as compared to the ferromagnet. At the boundaries the profiles show deviations from bulk behavior because of the broken exchange right at the boundary. b) the spin accumulation at the right end of the system as function of driving frequencyω0; as for the ferromagnet there are two regimes separated by a critical frequency ωcrit. magnet, the system reaches a steady state after a tran- sientphasecharacterizedbyalimitedexponentialgrowth on a time scale τt, which is roughly the same as for the ferromagnet. In the steady state the sublattice-resolved in-plane anglesA,Bϕboth follow exactly the same profile as the ferromagnet, i.e. eq. (9), but with a phase differ- ence ofπbetween the two sublattices because of the an- tiferromagnetic order (data for the antiferromagnet not shown). The spin accumulation deviates from the behavior of a ferromagnet as depicted in fig. 6 a). The bulk profiles (away from boundaries at x= 0andx=Nxa) are iden- tical in the two sublattices,ASz=BSz. Hence, a measur- able spin accumulation is present, but it is two orders of magnitude lower than in comparable ferromagnetic cases. This is the aforementioned exchange reduction. If we consider the spin accumulation Szin bulk, in the data in fig. 6 a) hardly a space dependence is observed in contrast to the ferromagnet, where Sl zhas a finite slope. The antiferromagnet exhibits this in the same way, but it is also much smaller and the profile becomes roughly constant. Contrary to the ferromagnet, there are distur- bances at the boundaries in the profile of Szwhich we already discussed before. Driving the antiferromagnet with the time-dependent boundary condition eq. (4) at frequency ω0leads to the very same two different regimes as for ferromagnets, a linear regime up to a critical frequency ωcritand above— in the nonlinear regime—phase slips occur. These phase slips reduce the winding number, lead to the excitation of spin waves, and a further increase of the spin accu- mulation is not possible. We quantify this behavior in a similar way as for the ferromagnet. It is, however, not possible to use the spin accumulation of the last spin SN zas a measure because of the deviating profile at the boundary. Instead, we take the spin accumulation at the end of the bulk in form of a spatial average over the spins in the range xl/a∈[4900,4920],Send z:=/angbracketleftbig Sl z/angbracketrightbig [4900,4920].This range is chosen such that it is sufficiently separated from the boundary. The data for the ω0dependence of thespinaccumulationareshowninfig.6, panelb): These show that critical frequencies takes roughly same values for ferro- and antiferromagnets, a result which has been tested and confirmed for another parameter set with dif- ferentNx,α, anddz. For the data set shown here the value isωcrittJ≈−5.75×10−4. However, the decrease of the spin accumulation Send zwith increasing driving fre- quencyω0in the nonlinear regime is less pronounced for antiferromagnets. We also calculated the ω0dependence of the time-averaged dissipated power /angbracketleftPdiss/angbracketrightand of the phase-slip rate Γps, both shown in fig. 4. Similar to other features these properties behave for the antiferromagnets very much as for ferromagnets: below ωcritthe dissipated power shows exactly the same dependence and above it is dominated by phase slips. However, a difference is that aboveωcritthe dissipated power increases faster with ω0. One reason for this might be the dynamics of spin waves that very much differ between ferro- and antiferromag- nets. The phase-slip rate differs slightly, however, this seems to be solely due to the fact that ωcritdiffers for ferro- and antiferromagnets. When Γpsis plotted versus ω0−ωcrit, both curves match almost. The next step is to consider finite temperature. Again this requires a finite cross section for which we use Nx×Ny×Nz= 2000×4×4and we test two temper- atures,kBT/J = 10−2andkBT/J = 10−4. As before, the magnetic response is very similar to that of a ferro- magnet: the in-plane angles follow the zero-temperature profiles, as well as does the average spin accumulation for the lower of the two temperatures. The only major difference is the ratio of the spin-superfluid spin accumu- lation to the thermal fluctuations, which is two orders of magnitude smaller as a result of the lower spin-superfluid signal and an equal strength of the fluctuations. For the higher temperature, this even leads to an average Szthat is essentially zero. This does not mean that there is no8 spin-superfluid spin accumulation, but rather that the available numerical data are not sufficient to resolve it and more averaging is needed. Note that the in-plane angle is not affected by this—it is as robust against the fluctuations just as for the ferromagnet. V. DISCUSSION AND CONCLUSION Our comparative study addresses spin superfluidity in ferro- and antiferromagnets, where one should bear in mind that the former are less promising for spin super- fluidity as the latter because of the negative influence of the stray field [22]. Nevertheless, the former can help to understand the behavior of the latter, which we utilize in this work. One of the striking features of spin super- fluidity is the transport range that leads to a spin ac- cumulation at the end of the system Sz(L)(see eqs. (9) and (10)) that does depend on the system length L— a completely different situation compared to spin-wave transport where the intensity decays exponentially with the distance. However, this non-exponential decay does not imply the possibility of an infinite transport range since with increasing system size the critical frequency lowers until no undisturbed spin superfluid is possible anymore. We present a full analytical solution for the steady state of the ferromagnet, which slightly deviates from the analytical theory reported before [19, 21]. This theory is tested numerically by the full atomistic model, which allows to test the robustness of the spin-superfluid trans- port against varying boundary conditions, against high excitation frequencies and finite temperature. We show that this kind of transport is remarkably robust: bound- ary conditions and also elevated temperature hardly hamper the magnets spin-superfluid response. Furthermore, we identify the critical frequency ωcrit— a manifestation of the Landau criterion—as the limiting factor for the range of this transport. Above this critical frequency phase slips occur, which also sets a limit to the spin accumulation that can be achieved. In ref. [38] another limitation on the spin current of such a spin su- perfluid is discussed, which rests on the fact that |Sz| is bounded above. But the estimated values would re- quire an out-of-plane component that takes quite large values|Sz|>0.1, which our simulations reveal to be hardly possible even for low damping. This is in particu-lar true for the case for antiferromagnets and, therefore, we conclude that the critical frequency—and therefore the phase slips—is a more relevant limitation on spin su- perfluid transport. The direct comparison of antiferromagnets to ferro- magnets shows that both exhibit the very same behavior: Driven by an in-plane rotation, both form an in-plane spin spiral that exhibits exactly the same behavior, in- cluding a spin accumulation in form of an out-of-plane magnetization. Antiferromagnets show in principle the same transport range as ferromagnets with a spin accu- mulation at the end of the system independent of the system length, provided the excitation frequency ω0is kept constant ( ω0itself depends on the magnets geome- try in experimental setups, see eq. (B12)). Furthermore, the critical frequency takes very similar value for the two types of magnets. This general accordance of spin super- fluidity for both types of magnets is in contrast to spin- wavetransportthatisknowntobedifferentforferro-and antiferromagnets[39]. Yetthereisamajordeviation: the antiferromagnetic exchange lowers tremendously the spin accumulation. Ourstudyalsocoversanexaminationofthedissipation ofaspinsuperfluidandoftheeffectoffinitetemperature. We proof the principle robustness of spin superfluidity against thermal fluctuations, i.e. that quite high temper- atures are required before thermal phase slips start to hamper the transport. But the fluctuations are a prob- lem from the numerical side as these require integration overalargeamountofdatainordertoidentifyanon-zero mean spin accumulation. The signal-to-noise ratio might be a problem in experimental setups as well and it could be more promising to measure rather the in-plane an- gleϕ, which is more robust against thermal fluctuations and which always delivers a clear signal in the cases we investigated here. A measurement of ϕcan be done in two ways: either by its time evolution, i.e. the preces- sion frequency ω0, or spatially resolved by measuring the formation of the in-plane spin spiral. ACKNOWLEDGMENTS Financial support by the Deutsche Forschungsgemein- schaft (DFG) via the SFB 767 “Controlled Nanosystems: Interaction and Interfacing to the Macroscale” and the program “Hematite: A new paradigm for antiferromag- netic spin transport” is gratefully acknowledged. [1] M. Wu and A. Hoffmann, eds., Recent Advances in Mag- netic Insulators – From Spintronics to Microwave Appli- cations, Solid State Physics (Academic Press, 2013). [2] K. Nakata, P. Simon, and D. Loss, Journal of Physics D: Applied Physics 50, 114004 (2017). [3] G. E. W. Bauer, E. Saitoh, and B. J. van Wees, Nature Materials 11, 391 (2012).[4] V. Cherepanov, I. Kolokolov, and V. L’vov, Physics Re- ports229, 81 (1993). [5] A. V. Chumak, A. A. Serga, and B. Hillebrands, Nature Communications 5, 4700 (2014). [6] A. V. Chumak, V. I. Vasyuchka, A. A. Serga, and B. Hillebrands, Nature Physics 11, 453 (2015). [7] S. O. Demokritov and A. N. Slavin, eds., Magnonics.9 From Fundamentals to Applications (Springer-Verlag, Berlin, Heidelberg, 2013). [8] U. Ritzmann, D. Hinzke, and U. Nowak, Phys. Rev. B 89, 024409 (2014). [9] Y. Zhang, T.-H. Chuang, K. Zakeri, and J. Kirschner, Phys. Rev. Lett. 109, 087203 (2012). [10] S. Hoffman, K. Sato, and Y. Tserkovnyak, Phys. Rev. B 88, 064408 (2013). [11] A. Kehlberger, U. Ritzmann, D. Hinzke, E.-J. Guo, J. Cramer, G. Jakob, M. C. Onbasli, D. H. Kim, C. A. Ross, M. B. Jungfleisch, B. Hillebrands, U. Nowak, and M. Kläui, Phys. Rev. Lett. 115, 096602 (2015). [12] L. J. Cornelissen, K. J. H. Peters, G. E. W. Bauer, R. A. Duine, and B. J. van Wees, Phys. Rev. B 94, 014412 (2016). [13] U. Ritzmann, D. Hinzke, and U. Nowak, Phys. Rev. B 95, 054411 (2017). [14] B. I. Halperin and P. C. Hohenberg, Phys. Rev. 188, 898 (1969). [15] E. B. Sonin, Zh. Eksp. Teor. Fiz. 74, 2097 (1978). [16] E. B. Sonin, Phys. Rev. B 95, 144432 (2017). [17] Y.M.Bunkov,“Spinsupercurrentandnovelpropertiesof NMR in3He,” inProgress in Low Temperature Physics , Vol. 14, edited by W. P. Halperin (Elsevier, 1995) pp. 69–158. [18] J. König, M. C. Bønsager, and A. H. MacDonald, Phys. Rev. Lett. 87, 187202 (2001). [19] S. Takei and Y. Tserkovnyak, Phys. Rev. Lett. 112, 227201 (2014). [20] S. Takei, B. I. Halperin, A. Yacoby, and Y. Tserkovnyak, Phys. Rev. B 90, 094408 (2014). [21] B. Flebus, S. A. Bender, Y. Tserkovnyak, and R. A. Duine, Phys. Rev. Lett. 116, 117201 (2016). [22] H. Skarsvåg, C. Holmqvist, and A. Brataas, Phys. Rev. Lett.115, 237201 (2015). [23] W. Yuan, Q. Zhu, T. Su, Y. Yao, W. Xing, Y. Chen, Y. Ma, X. Lin, J. Shi, R. Shin- dou, X. C. Xie, and W. Han, Science Ad- vances4, eaat1098 (2018), 10.1126/sciadv.aat1098, http://advances.sciencemag.org/content/4/4/eaat1098.full.pdf. [24] R.Mondal, M.Berritta, andP.M.Oppeneer,Phys.Rev. B94, 144419 (2016). [25] D. A. Bozhko, A. A. Serga, P. Clausen, V. I. Vasyuchka, F. Heussner, G. A. Melkov, A. Pomyalov, V. S. L’vov, and B. Hillebrands, Nature Physics 12, 1057 (2016). [26] E. B. Sonin, “Comment on "supercurrent in a room tem- perature Bose-Einstein magnon condensate",” (2016), arXiv:1607.04720. [27] D. A. Bozhko, A. A. Serga, P. Clausen, V. I. Vasyuchka, G. A. Melkov, V. S. L’vov, and B. Hillebrands, “On supercurrents in Bose-Einstein magnon condensates in YIG ferrimagnet,” (2016), arXiv:1608.01813. [28] R. Lebrun, A. Ross, S. A. Bender, A. Qaiumzadeh, L. Baldrati, J. Cramer, A. Brataas, R. A. Duine, and M. Kläui, Nature 561, 222 (2018). [29] L. J. Cornelissen, J. Liu, R. A. Duine, J. Ben Youssef, and B. J. van Wees, Nature Physics 11, 1022 (2015). [30] U. Nowak, “Classical spin models,” in Handbook of Mag- netism and Advanced Magnetic Materials , Vol. 2, edited by H. Kronmüller and S. Parkin (John Wiley & Sons, Ltd, 2007) available under http://nbn-resolving.de/ urn:nbn:de:bsz:352-opus-92390 . [31] L. Landau and E. Lifshitz, Phys. Z. Sowjetunion 8, 153 (1935).[32] T. L. Gilbert, Physical Review 100, 1243 (1955). [33] T. L. Gilbert, IEEE Transactions on Magnetics 40, 3443 (2004). [34] W. F. Brown Jr., Phys. Rev. 130, 1677 (1963). [35] M. P. Magiera, L. Brendel, D. E. Wolf, and U. Nowak, Europhysics Letters 87, 26002 (2009). [36] E. B. Sonin, Advances in Physics 59, 181 (2010), https://doi.org/10.1080/00018731003739943. [37] H. Ochoa, R. Zarzuela, and Y. Tserkovnyak, Phys. Rev. B98, 054424 (2018). [38] Y. Tserkovnyak and M. Kläui, Phys. Rev. Lett. 119, 187705 (2017). [39] F. Keffer, H. Kaplan, and Y. Yafet, Amer- ican Journal of Physics 21, 250 (1953), https://doi.org/10.1119/1.1933416. [40] A. Manchon, J. Železný, I. M. Miron, T. Jungwirth, J. Sinova, A. Thiaville, K. Garello, and P. Gambardella, Rev. Mod. Phys. 91, 035004 (2019). Appendix A: Analytical theory for a 1D ferromagnet The ferromagnet in the micromagnetic approximation under the assumption of small out-of-plane component, |Sz|/lessmuch1, is described by the LLG equation in cylindrical coordinates, eqs. (7) and (8). Assuming translational invariancealong y-andzdirectionleadstoa1Dproblem: µS γ˙ϕ=Ja2∂2 xSz+Ja2Sz(∂xϕ)2+ 2dzSz−αµS γ˙Sz (A1) µS γ˙Sz=−Ja2∂2 xϕ+αµS γ˙ϕ. (A2) Steady state means ˙ϕ=ω0and ˙Sz= 0. This allows to integrate the latter equation, sϕ(x,t) =α 2µSω0 γJ/parenleftbiggx−L a/parenrightbigg2 +ω0t+ϕ0,(A3) where the first integration constant follows from the Neu- mann boundary condition at the right end, ∂xϕ(L) = 0 (no outflow of spin current), and the second one satisfies the condition ˙ϕ=ω0and allows for an arbitrary phase ϕ0. This is inserted in the first equation, which then reads −Ja2∂2 xSz=−µSω0 γ+µ2 Sω2 0 γ2J/parenleftbigg αx−L a/parenrightbigg2 Sz+ 2dzSz. (A4) We argue that the second-derivative term can be ne- glected−Ja2∂2 xSz≈0. This is justified in a twofold manner: first we compared the relevance of all terms in that equation numerically by calculating those three terms from simulations of the full atomistic spin model, eq. (2). Indeed the result is that in steady state the second-derivative term is several orders of magnitude smaller compared to the other two. The second reason follows a-posteriori from the calculated solution and is10 spin injector (using SHE) spins not subjected to SHEspins driven by SHEspins not subjected to SHE Figure 7. 1D setup for calculation of the excitation frequency ω0of a magnet driven by a spin injector utilizing the spin-Hall effecttoexertexternaltorquesonthespins. Thesetorquesare applied in the region [l1,l2]and vanish outside. Furthermore, the Gilbert damping in [l1,l2]is enhanced by αd. The ground stateSis in-plane, the spin accumulation µperpendicular. explainedbelow. From −Ja2∂2 xSz≈0followsthesteady- state solution for Sz: sSz=µSω0 2γdz 1 +µ2 Sω2 0 2γ2Jdz/parenleftbig αx−L a/parenrightbig2. (A5) This solution does not fulfill eq. (A4), however, we can insert it and calculate the deviation by calculating ∂2 xsSz=−2µSω0 γJα2 a2sS2 z+ 4/parenleftbiggµSω0 γJ/parenrightbigg2α4(x−L)2 a4sS3 z =O/parenleftbig S2 z/parenrightbig . This allows the conclusion that the correction by taking the second derivative into account is of higher order in Szand neglecting this is consistent with the original as- sumption|Sz|/lessmuch1. Hence, eqs. (A3) and (A5) form the analytical solution for a 1D setup. Appendix B: Frequency of a spin superfluid The usual excitation of a spin current in a magnet rests on a spin accumulation µat an interface between the magnet and a heavy metal, which is created by an electrical current. Normally for that the spin-Hall effect is utilized. The aim of this appendix is to calculate the resulting excitation frequency ω0of a spin superfluid. We assume here that the spin accumulation is per- pendicular to the magnets ground state, i.e. µ∝ez. Consequently, there is an additional damping-like torque [22, 40] in the LLG equation (here written as viscousdamping): ˙Sl=−γ µSSl×Hl+αlSl×˙Sl+α/prime lSl×/parenleftbigg Sl×µl ~/parenrightbigg . (B1) A subsetVdof the total volume of the magnet is driven, i.e. subjected to the additional torques and the driving also creates an enhanced damping α/prime lwithinVd: µl=/braceleftbigg µdezforrl∈Vd 0else(B2) αl=α0+α/prime lwithα/prime l=/braceleftbigg αdforrl∈Vd 0else.(B3) α0is the intrinsic Gilbert damping of the magnet. To proceed we consider the LLG equation in the fol- lowing form, resolved for the time derivative: ˙Sl=−γ µS(1 +α2 l)Sl×/parenleftbig Hl+αlSl×Hl/parenrightbig +Tl 1Sl×Al+Tl 2Sl×/parenleftbig Sl×Al/parenrightbig .(B4) Tl 1andTl 2parameterize arbitrary additional torques with respecttoanaxis Alandforthespecificchoice Al=µl/~, Tl 1=αlα/prime l/(1+α2 l)andTl 2=−α/prime l/(1+α2 l)eq. (B4) is equiva- lent to eq. (B1). However, for the sake of generality we consider for the calculation eq. (B4). Assuming Al∝ez and using cylindrical coordinates and again the micro- magnetic approximation, this form of the LLG reads µS γ˙ϕ=Ja2Sz|∇ϕ|2+ 2dzSz−αµS γ˙Sz −µS γAz(T1+αT2) (B5) µS γ˙Sz=−Ja2∆ϕ+αµS γ˙ϕ+µS γAz(αT1−T2),(B6) an extension of eqs. (7) and (8). In the same spirit as in appendix A we can solve these equations in one di- mension in steady-state (assuming ˙Sz= 0and ˙ϕ=ω0), where the geometry depicted in fig. 7 is assumed. We ap- ply the external spin accumulation in the interval [l1,l2], whereas the total magnet expands over [0,L]. Therefore, T1,2(x) =/braceleftbigg Td 1,2forx∈[l1,l2] 0else A(x) =/braceleftbigg Ad zezforx∈[l1,l2] 0else.11 In the 1D setup eq. (B6) reads ∂2 xϕ=α(x)µS γJa2ω0+µS γJa2Az(x) [α(x)T1(x)−T2(x)] = =:¯ω0/bracehtipdownleft/bracehtipupright/bracehtipupleft/bracehtipdownright α0µS γJa2ω0 forx∈[0,l1] (α0+αd)µS γJa2ω0 /bracehtipupleft/bracehtipdownright/bracehtipdownleft/bracehtipupright =:¯ω/prime 0+µS γJa2Ad z/bracketleftbig (α0+αd)Td 1−Td 2/bracketrightbig /bracehtipupleft/bracehtipdownright/bracehtipdownleft /bracehtipupright =:tforx∈[l1,l2] α0µS γJa2ω0 forx∈[l2,L], (B7) which can be integrated. There are six boundary conditions to consider, each one at the left and right end of the magnet, where we assume a Neumann condition ∂xϕ(0) =∂xϕ(L) = 0, i.e. no outflow of spin currents. Furthermore, ϕ and∂xϕmust be continuous at l1andl2, delivering four internal boundary conditions. But there is another condition, a gauge condition for ϕ, which allows to add an arbitrary constant phase to ϕ(x)without altering the physics. (In practice this gauge phase depends on the prehistory of the magnet, i.e. on how it had reached its steady state, and also which exact instant in time is considered.) As gauge we use ϕ(0) = 0. Altogether there are 6 integration constants and the unknown frequency ω0in combination with 6 boundary conditions and a gauge, such that the problem has a unique solution. As result we obtain ϕ= 1 2¯ω0x2forx∈[0,l1] 1 2(¯ω/prime 0+t)x2+ (¯ω0−¯ω/prime 0−t)l1x+1 2(¯ω/prime 0−¯ω0+t)l2 1forx∈[l1,l2] 1 2¯ω0x2+ (¯ω/prime 0−¯ω0+t)/bracketleftbig (l2−l1)x+1 2(l2 1−l2 2)/bracketrightbig forx∈[l2,L](B8) Sz=µS γω0+Az(x) [T1(x) +α(x)T2(x)] Ja2(∂xϕ)2+ 2dz(B9) and, importantly, we also gain ω0=−Ad z/bracketleftbig (α0+αd)Td 1−Td 2/bracketrightbig (l2−l1) α0L+αd(l2−l1). (B10) This holds true for arbitrary torques taking form eq. (B4). If the specific case of the spin injector utiliz- ing the spin-Hall effect is considered, then inserting the parameters T1,T2andAreads ω0=−µd ~αd /bracehtipupleft/bracehtipdownright/bracehtipdownleft/bracehtipupright =:τ·l2−l1 α0L+αd(l2−l1).(B11) The former factor τis the strength of the spin-Hall effect on the magnet [40]: τ=γ Ms~ 2eηϑSHjel1 d, with spin transparency of the interface η, spin-Hall an- gleϑSH, saturation magnetization Msand thickness d of the magnet. jelis the electric current density. The latter factor in eq. (B11) is a geometric factor that is ba- sically the ratio between the driven volume l2−l1and the total volume L, weighted with the total damping of the magnet, where the Gilbert-damping enhancement can beexpressed as [22] αd=g⊥~2 2e2γ Msd, with transverse spin mixing conductance g⊥of the in- terface. This rigorous derivation holds only true for 1D ferromagnets, however, the natural extension to 2D and 3D is given by ω0=−τ·Vd α0V+αdVd, (B12) whereVis the magnets total and Vdthe driven volume. The validity of this expression has been checked numeri- cally for 1D and 2D systems using various geometries by investigating the full atomistic LLG eq. (B4). As a result we obtain very good agreement with the analytical calcu- lation except for two cases. First, when the assumption |Sz|/lessmuch 1is violated and second if the setup is not ef- fectively one dimensional, i.e. if the system is not driven over the entire transverse width. However, such a mis- match in usually small for realistic experimental setups. We furthermore did not only simulate ferromagnets, but12 also antiferromagnets with same parameters except for the sign of J. These simulations result in exactly the same frequencies ω0as the corresponding ferromagnetsand thus eqs. (B10) to (B12) are also valid for antiferro- magnets, even though note that the resulting spin accu- mulation deviates. | 2019-11-28 | We present a study on spin-superfluid transport based on an atomistic,
classical spin model. Easy-plane ferro- as well as antiferromagnets are
considered, which allows for a direct comparison of these two material classes
based on the same model assumptions. We find a spin-superfluid transport which
is robust against variations of the boundary conditions, thermal fluctuations,
and dissipation modeled via Gilbert damping. Though the spin accumulations is
smaller for antiferromagnets the range of the spin-superfluid transport turns
out to be identical for ferro- and antiferromagnets. Finally, we calculate and
explore the role of the driving frequency and especially the critical
frequency, where phase slips occur and the spin accumulation breaks down. | Transport properties of spin superfluids: comparing easy-plane ferro- and antiferromagnets | 1911.12786v1 |
Determination of the damping coecient of electrons in optically transparent glasses at the true resonance frequency in the ultraviolet from an analysis of the Lorentz-Maxwell model of dispersion Surajit Chakrabarti (Ramakrishna Mission Vidyamandira) Howrah, India The Lorentz-Maxwell model of dispersion of light has been analyzed in this paper to determine the true resonance frequency in the ultraviolet for the electrons in optically transparent glasses and the damping coecient at this frequency. For this we needed the refractive indices of glass in the optical frequency range. We argue that the true resonance condition in the absorption region prevails when the frequency at which the absorption coecient is maximum is the same as the frequency at which the average energy per cycle of the electrons is also a max- imum. We have simultaneously solved the two equations obtained from the two maxima conditions numerically to arrive at a unique solution for the true resonance frequency and the damping coecient at this frequency. Assuming the damping coecient to be constant over a small frequency range in the absorption region, we have determined the frequencies at which the extinction coecient and the re-
ectance are maxima. These frequencies match very well with the published data for silica glasses available from the literature. 1arXiv:1907.04499v1 [physics.optics] 10 Jul 20191 Introduction The Lorentz-Maxwell model of dispersion of electromagnetic waves in matter is very successful in describing the properties of matter under the action of electro- magnetic waves over its whole spectrum where the wavelength is large compared to the interatomic distances. The model is generally studied in the optical frequency range where only the oscillation of electrons bound to atoms and molecules is rel- evant for the study of dispersion. Two important parameters of the model namely the natural oscillation frequency and the plasma frequency of the electrons in a dielectric medium like glass can be easily determined from the refractive indices of a glass prism measured in the optical band [1] where glass is transparent. In a condensed system like glass one has to include the eect of the local eld on the electrons apart from the eld of the incident wave. This leads to another fre- quency which is conventionally known as the resonance frequency and is related to the plasma and the natural oscillation frequencies of the electron [2]. Though it is called the resonance frequency, there is no proof that the absorption coecient is maximum at this frequency. In order to study the absorption of EM waves in matter, a phenomenologi- cal variable called the damping coecient is introduced in the Lorentz-Maxwell model. Glass is opaque in the ultraviolet indicating that it has a strong absorption there. In scientic literature, there are innumerable experimental works which have studied the interaction of silica glasses with electromagnetic waves over its whole spectrum. A summary of these works can be found in Kitamura et al. [3]. From the experimental data on the extinction coecient for silica glass in the ultraviolet, we can nd the frequency at which this coecient is maximum. However, as far as we are aware, there has been no theoretical study so far which has determined this frequency by an analysis of the Lorentz-Maxwell model of dispersion. The main problem with the theoretical analysis is the fact that it has not been possible so far to determine the value of the damping coecient theoretically. 2In this work we have determined the damping coecient at the true resonance frequency which we dene to be the frequency at which the absorption coecient for the energy of the electromagnetic eld in the medium is maximum. We have done this theoretically by taking the natural oscillation frequency and the plasma frequency determined from the refractive indices of glass in the optical region as two known parameters of the Lorentz-Maxwell model. We have formed two algebraic equations containing the true resonance frequency and the damping coecient as two unknown variables. We have solved these two equations simultaneously by numerical method to nd a unique solution for the two variables. With the value of the damping coecient known, we have explored the anomalous dispersion region in the ultraviolet for glass. It is well known that the Kramers-Kronig relations [3] allow us to determine the imaginary part of the dielectric constant from an integral of the real part over the whole range of frequencies and vice versa. The theory is based on a very general causality argument and a linear response of the medium to an external perturbation. We have, on the other hand, determined the damping coecient of the Lorentz-Maxwell model of dispersion starting from the refractive indices in the optical region corresponding to the real value of the dielectric constant. From this we have extracted the information about the absorptive region in the ultraviolet corresponding to the imaginary part of the dielectric constant. In section 2 we give the outline of the Lorentz-Maxwell model. In section 3 we oer our physical argument for the method adopted to determine the damping coecient and the true resonance frequency. The next four sections are just an execution of these ideas. We conclude with a summary of the work. 32 Lorentz-Maxwell model of dispersion In the Lorentz model [4] of dispersion of light in a dense medium like a solid or liquid, electrons execute forced simple harmonic oscillations with damping in the combined eld of the incident electromagnetic wave of frequency !and the local eld. The local eld arises as a result of the interaction of the electron with the elds of other atoms close by. Without any loss of generality we can assume that the direction in which the electron is oscillating is the ydirection. We can write the equation of motion as y+
_y+!2 0y=qE0 0 me i!t: (1) whereE0 0is the amplitude of the eective eld acting on the electrons. Here !0is the natural oscillation frequency and
is the damping coecient of the electron. In the steady state the electron will oscillate at a frequency !of the incident wave though shifted in phase. E0 0is related to the amplitude of the eld ( Ei0) outside from where it is incident on the medium as E0 0=1 + 3 1 +DEi0: (2) The 3term in equation (2) arises as a result of the eect of local eld in the Lorentz-Lorenz theory of dielectric polarizability valid for an isotropic medium [5] whereis the electric susceptibility. Dis the depolarisation factor, a dimensionless number of the order of unity [6]. The dielectric function of the medium is given by = 1 +: (3) Using Maxwell's phenomenological relation =n2 cwherencis the complex refrac- tive index and the Lorentz-Lorenz equation [5], we arrive at the following equation for a number of resonance regions [7,8]. n2 c 1 n2 c+ 2=Nq2 30mX jfj !2 0j !2 i
j!: (4) 4Herefjis the fraction of electrons that have a natural oscillation frequency !0j and damping constant
jwith fj= 1.Nis the density of electrons taking part in dispersion. It is a common practice to assume a single dominant absorption frequency which is true in many practical cases and which makes the analysis simpler [9]. With this assumption fj= 1 and equation (4) can be written as n2 c= 1 +!2 p !2 n !2 i
!(5) where the plasma frequency !pis given by !2 p=Nq2 0m(6) and we dene !2 n=!2 0 !2 p 3: (7) In scientic literature [2], !0is known as the natural oscillation frequency of the electrons and !nis known conventionally as the resonance frequency. So far, au- thors have used some chosen values of the damping coecient
and the plasma frequency which mimic the absorptive properties of dielectric materials, in order to carry out model analysis [9]. We have actually determined the damping coe- cient from a prior knowledge of the natural oscillation frequency and the plasma frequency of a glass medium. In the optical limit where the absorption in glass is negligible we take
= 0. In this limit the refractive index is real and equation (5) reduces to n2= 1 +!2 p !2 n !2: (8) which is essentially the Sellmeier's formula [7] for dispersion in the frequency do- main with one absorption band. If we have a set of measurements of refractive indices of a glass prism for several optical wavelengths, we can determine !nand !pusing equation (8) [1].The resonance wavelength which falls in the ultraviolet region, has been determined in a similar work [7]. Once !nand!pare known, !0 can be determined using equation (7). 5In the absorptive region the dielectric function picks up an imaginary part given by n2 c==1+i2: (9) The refractive index ( n) and the extinction coecient ( ) known as optical con- stants are written as nc=n+i (10) whererepresents the attenuation factor of the amplitude of the electromagnetic wave in an absorptive medium. Using the last two equations and equation (5) we obtain for the real and imaginary parts of the complex dielectric function [10], 1=n2 2= 1 +!2 p(!2 n !2) (!2 n !2)2+
2!2(11) and 2= 2n=!2 p
! (!2 n !2)2+
2!2: (12) We can express n2and2as functions of frequency !using equations (11) and (12). The details and the nal expressions have been shown in Appendix A. The absorption coecient of the incident EM wave is given by [7,10], =2! c (13) wherecis the speed of light in vacuum. gives the attenuation coecient of the intensity of the incident wave. Intensity is the rate of
ow of energy per unit area normal to a surface. will be a maximum at the frequency at which the absorption of energy by the electrons from the EM eld is maximum. This gives the condition of resonance. It is a general practice to consider !ndened in equation (7) as the resonance frequency though there is no proof that the energy absorption is maximum at this frequency. So we do not assume a priori !nto be the resonance frequency. In the next section we will describe our strategy to nd the true resonance frequency and in the results section we will see that the true resonance frequency is dierent from both !0and!nand lies between them. There 6is no real reason to call !nthe resonance frequency. We treat the true resonance frequency as an unknown variable to be found from our analysis. The damping coecient
is introduced in the Lorentz model to explain absorp- tion. We model
such that it is zero in the optical band and upto the frequency !n. In the absorption band we assume that
is constant from frequency !nto !0. Above!0,
falls down and rises again to another constant value of
in the next resonance region if the material under study has one. With this model for
in mind we can extrapolate equation (8) to nd !nand!p. In the next section we will explain how to get the constant value of
in the absorption region and the true resonance frequency. Even if the system under study may have several absorption bands, we can study it with the assumption of a single resonance region. The optical waves os- cillate the outermost electrons of atoms and molecules having the lowest natural frequencies and as a result we get the phenomenon of refraction. With an anal- ysis of the refractive indices in the optical region under this assumption of single resonance, we are most likely to nd information about the absorption band with the lowest natural oscillation frequency in the ultraviolet closest to the optical band. This will of course depend on the strength of the resonance. The justica- tion of the single resonance calculations with the chosen model for
can be found from the results of our theoretical calculations which will be found to match the experimental results very well. 73 Physical argument for the method adopted to determine the damping coecient at the true resonance frequency From various experiments on the absorption of EM waves in matter, we know that the absorption coecient ( ) attains a maximum value at a characteristic frequency. We try to nd this frequency where is maximum. We dierentiate with respect to frequency and equating the derivative to zero get one equation. However, we have two unknown variables in the theory - the damping coecient and the true resonance frequency. We look for a second equation. The incident EM wave interacts with the electrons bound to the atoms and molecules.The electrons execute a forced simple harmonic oscillation with damp- ing. The total energy of the electron is time dependent, as the electron is being perturbed by a time dependent harmonic force.The average energy of the electron per cycle can be worked out easily [11]. We nd the frequency at which this av- erage energy per cycle is maximum. This leads to another equation involving the two unknown variables. When the frequency at which is maximum is the same as the frequency at which the average energy per cycle of the electron is also a maximum, the electromagnetic wave will share its energy most with the electrons and will be attenuated most. This will constitute the true condition of resonance. By solving the two equations simultaneously using numerical method, we nd both the variables. We call the characteristic frequency, the true resonance frequency !tand the damping coecient at the true resonance frequency
t. Heitler [12] has proposed a quantum theory of the phenomenon of damping. According to this theory the damping coecient is dependent on frequency though of a very slowly varying nature near resonances. This gives support to our earlier assumption that the damping coecient is a constant within a small frequency range about the resonance frequency. However, it can be taken as zero in the 8optical band where glass is transparent and absorption is negligible. 4 Condition for the maximum of the absorption coecient as a function of frequency Our aim in this section is to nd the frequency at which is maximum. We rst dierentiate with respect to !assuming
constant. In order to nd the derivative of we rst dierentiate equations (11) and (12) with respect to !. We nd two algebraic equations involvingdn d!andd d!. By eliminatingdn d!from the two equations, we get the expression ford d!and henced d!using equation (13). We have shown the dierentiations in Appendix B. Eliminatingdn d!between equations (B.2) and (B.3) we get 2d d!(n+2 n) =A B C(14) where A=!2 p(!2 n !2)2[
2 n!] (15) and B=!2 p!
[!
2 4(!2 n !2)! 2 n!2 n
] (16) and C= [(!2 n !2)2+
2!2]2: (17) From this we get d d!= c[2 +n ! n2+2A B C]: (18) Ifis maximum thend d!should be zero. So we write at the maximum !(A B) C= 2 n(n2+2): (19) It is to be noted that two sides of equation (19) are dimensionless and they will be compared later numerically to nd the solution for the true resonance frequency and the damping coecient. 95 Condition for the maximum of the average en- ergy per cycle of the electron as a function of frequency In the steady state the electron will oscillate at a frequency !as given by the steady state solution of equation (1) and the total energy of the system averaged over a period is given by [11], E(!) =1 4(qE0 0)2 m(!2+!2 0) [(!2 0 !2)2+ (!
)2]=1 4(qE0 0)2 mg(!) (20) where g(!) =(!2+!2 0) [(!2 0 !2)2+ (!
)2]: (21) Equation (2) shows the relationship between the incident electric eld and the eld acting on an electron. With the variation of frequency in the ultraviolet we can imagine that the amplitude of the incident eld is kept constant. However, the amplitudeE0 0is dependent on which is frequency dependent. Lorentz theory is based on the assumption that the response of the medium to the external eld is small [13]. In equation (2), appears both in the numerator as well as in the denominator. With the depolarization factor Dpositive, any variation of in the numerator will be oset to some extent by the variation in in the denominator. So we neglect the variation of the term E0 0with frequency and assume it to be constant. To nd the derivative of the average energy per cycle E(!), it is sucient to nd the derivative of the function g(!) given by equation (21) with respect to !. Equating the derivative to zero, we nd the condition at which the average energy per cycle is maximum. It turns out that the frequency is given by !=!0[r 4 (
!0)2 1]1 2: (22) If the incident electromagnetic wave can oscillate the bound electrons steadily at frequency!given by the last equation, then the wave has to deliver maximum energy per cycle and its absorption will be maximum. 10It is clear from equation (22) that for real values of !we should have the ratio f=
!0<p 3. By trial we take several positive values of fupto its maximum of p 3. For each value of fwe nd the values of
and!using equation (22) with the known value of !0. With these values we determine the refractive index nand the extinction coecient in the resonance region using equations (A.3) and (A.4) respectively. We put these values in equation (19) and try to see for which value offthis equation is satised. From fwhich satises equation (19) we calculate
t and!tusing equation (22). 6 Determination of the damping coecient and the true resonance frequency The results of an experiment performed with a prism made of
int glass have been reported by Chakrabarti [1]. In this experiment1 n2 1has been plotted against the inverse wavelength squared at optical range following equation (8). From this plot we have determined the values of !n,!p:The value of !0has been estimated using equation (7). The errors in these frequencies are less than 1% :Refractive indices as a function of wavelengths have been shown in table 1 and the parameters needed for this work have been shown in table 2. It has been shown in the discussion following equation (22) that the maximum value off=
!0isp 3. So we take trial values of flike 0.1, 0.2 upto 1.7. For all these values offwe determine
and then!using equation (22). We then determine n andusing equations (A.3) and (A.4) respectively and calculate A,B,Caccording to equations (15),(16),(17) respectively. With these values we try to see whether they satisfy equation (19) or not. The algorithm for this is as follows: suppose we call the left side of equation (19), Y1 and the right side, Y2. Now we form a parameter, Y=Y1 Y2. For all values of ffrom 0.1 to 1.7 we calculate Y and nd between which two values of f, the sign of Ychanges. In our case there 11was only one sign change in Ybetweenfvalues 0.6 and 0.7. Now we check this region more closely, that is between 0.60 to 0.70 at an interval of 0.01. We nd that forf= 0:65,Y= 0:001 and for f= 0:66,Y= 0:053. So there is a zero crossing between fvalues 0.65 and 0.66. Proceeding similarly we nally nd that forf= 0:6501,Y= 0:0005 and for f= 0:6502,Y= 0:000055. So the solution lies between 0.6501 and 0.6502. We take the solution as f= 0:65015 with an error of 0:00005 which is just 0:008%. From this value of fwe get
tsince we know the value of !0. We get the value of !tby putting the value of fin equation (22). These values are shown in table 3. The values of the damping coecient and the true resonance frequency turn out to be
t= 11:61015rad/s with an error of only 0 :76% and !t= 16:81015rad/s with an error similar to the error in !0.!tlies between !nand!0. This solution is unique. We nd that !tdiers signicantly from !n. We have presented the parameters at true resonance frequency in table 4. The maximum value of the absorption coecientat the true resonance frequency has come out to be 8 :18107m 1 corresponding to an attenuation length 12 :2 nm. Once we determine the damping coecient, we can determine the absorption coecient in a small frequency range about !twhere we assume
is constant and equal to
t. In gure 1 we plot as a function of x=! !0betweenx= !n !0= 0:81 to 1. We assume that within the frequency range !nto!0,
remains constant. We clearly nd that attains a maximum value at x=!t !0= 0:94: Jackson [14] has given the frequency range for absorption coecient in the ultraviolet as well as the plasma frequency of water. In the ultraviolet, we are concerned with electronic oscillations as the most important component responsible 12for dispersion. So the properties of glass will not be too dierent from water at these frequencies. This work [14] shows that the maximum value of is around 108m 1which is approximately the same as the maximum value that we have obtained. The frequency at which the maximum occurs is also the same in order of magnitude as !t:The full width at half maximum of the absorption curve can be read o approximately. It is determined to be approximately 15 1015rad/s which is of the same order of magnitude that we have obtained for the value of
t. The damping coecient determined in this work is rather large.This broad ab- sorption in the ultraviolet is due to the outer electrons in the atoms and molecules of the solid which take part in the process of dispersion. The outer electrons are aected by the collisions and the electric elds of the neighbouring atoms. Con- sequently an extensive region of continuous absorption is obtained in solids and liquids [8,15]. So this large value of
tis expected. This value of
tis valid only in the resonance region. 7 Some other results We assume the damping coecient
tis constant within the small frequency range from!nto!0in which!tlies. We now nd the frequency !at which the extinction coecientis maximum, assuming !is close to!t. Forto be maximum,d d! should be zero and from equation (14) this should occur when A=B:Here we scan a parameter! !0from 0.1 to 1.0. For each value of !and known
twe calculate nandusing equations (A.3) and (A.4) respectively. We then check whether the condition A=Bis satised. Once again we nd a unique solution with ! !0= 0:84843:This corresponds to an angular frequency != 15:11015rad/s. From this we get = 2:401015Hz 13corresponding to a wavelength = 0:125m: In gure 2 we have plotted determined by our theoretical analysis, as a function of ! !0:This gure clearly shows that has attained a maximum value. The maximum value ofthat we have obtained at !=!is 0.769. This is of the same order of magnitude as the maximum value shown in gure 2 of Kitamura et al.[3] which gives the maximum of in the ultraviolet at a wavelength very near to 0 :12m. This is very close to the wavelength that we have obtained. In the last column of table 3 we show the value of !.that we have determined may be seen from gure 1 of [3] to be the nearest to the optical wavelengths on long wavelength side as we claimed in last paragraph of section 2. The data in gure 1 of [3] may be showing other resonance/s at shorter wavelengths or larger frequencies. In gure 3 we show a plot of re
ectance Ras a function of frequency in the absorption region where Ris the normal re
ectivity dened in [3] as R=(n 1)2+2 (n+ 1)2+2: (23) We nd that this function has a peak at! !0= 0:908 with the maximum value of R= 0:118. The peak position corresponds to != 16:21015rad/s:The re
ectance peak position corresponds to an energy 10.6 eV. Experimentally, silica glasses show the lowest frequency re
ectance peak in the ultraviolet region at 10.2 eV [16] which is very close to what we have found theoretically. This is once again as we claimed in the last paragraph of section 2. Sigel has shown in gure 3 [16] that the re
ectance peak position is the same for two glassy materials though their re
ectance values at this frequency are dierent. Similarly the actual values of refractive indices and extinction coecients at the same frequency can vary signicantly due to glass manufacturing process [3]. Data for these coecients in the absorption region for
int glasses were not available in the literature. However, the order of magnitude of these coecients is similar to that of other silica glasses. Since we are getting 14the peak positions of the extinction coecient and the re
ectance data very close to the published data for silica glasses, we conclude that the value of the damping coecient obtained by us is correct. We have shown in table 5 the values of n,,, the absorption length1 and g(!) at frequencies close to !t. They have been calculated at three frequencies !0, !and!nfor the same
t. Comparing tables 4 and 5 we nd that andg(!) are indeed maxima at !t. The anomalous nature of variation of the refractive index is evident from the values of nat frequencies around !t. In section 4 we determined the condition for to be maximum. In a similar way we can nd the condition for the refractive index nto be an extremum by equating its derivative to zero. This derivative can be obtained by eliminatingd d! from equations (B.2) and (B.3). Interestingly, this condition gives two solutions for the frequency. The refractive index is maximum at one and minimum at the other frequency. We nd that the Lorentz-Maxwell model of dispersion reproduces all the features of anomalous dispersion in the absorption region as observed in actual experiments [3]. 8 Conclusions The Lorentz-Maxwell model of dispersion of electromagnetic waves in matter has been studied in this paper with an analysis of the phenomenon of absorption in the ultraviolet in dielectrics like
int glass. We have shown that if we know the refractive indices of glass fairly accurately in the optical frequencies, we can explore the anomalous dispersion region in the ultraviolet quantitatively. The key nding of this work is that the damping coecient of the model can be determined by a simple argument. We also determine the frequency at which the absorption coecient is maximum. We call this the true resonance frequency. In the optical region where glass is transparent, the damping coecient can be assumed to be 15zero. In the absorptive part the damping coecient has been taken to be a constant within a short range of frequencies. The value of the damping coecient matches in order of magnitude with the experimental width of the absorption coecient data for water available from literature. Once the damping coecient is determined, we can nd the frequencies at which the extinction coecient and the re
ectance are maxima. These frequencies match very well with the experimental data available in the literature for silica glasses. This indirectly shows that the value of
determined by us is correct. Our assumption of a single resonance should give us the information of the absorption band closest to the optical frequencies. We actually observe this by comparing the peak positions of the extinction coecient and re
ectance data obtained by us with that found from literature. Refractive indices estimated at dierent frequencies close to the true resonance frequency in the absorption region reveal the anomalous nature of dispersion. All the features of dispersion by a dielectric like glass in the ultraviolet absorption region have been reproduced from our theoretical analysis of the Lorentz-Maxwell model. 9 Acknowledgment The author would like to thank Prof. Jayanta Kumar Bhattacharjee for some helpful discussions. Thanks are due to Prof. Debashis Mukherjee for making some helpful comments on the paper. 16Appendix A Equations (11) and (12) can be easily inverted and we get [2], n2=1 2[(2 1+2 2)1 2+1] (A.1) and 2=1 2[(2 1+2 2)1 2 1]: (A.2) We express 1and2as functions of frequency using equations (11) and (12) respectively. After a fairly straightforward algebra we arrive at the nal expressions forn2and2as functions of frequency !. n2=[!4 p
2!2+ (!04+
2!2+!2 p!02)2]1 2+ (!04+
2!2+!2 p!02) 2(!04+
2!2)(A.3) 2=[!4 p
2!2+ (!04+
2!2+!2 p!02)2]1 2 (!04+
2!2+!2 p!02) 2(!04+
2!2)(A.4) where !02=!2 n !2: (A.5) These equations are exact and will be used for determining nandin the resonance region in the ultraviolet. It can be easily checked that in the limit
tending to zero, becomes zero at all frequencies and vice versa. Thus the medium is transparent in the optical frequencies as it should. In this limit the refractive index nsatises a relation which has been used in the rst place to get the parameters !nand!p[1]. Appendix B We rst nd the derivative of1 (!2n !2)2+
2!2with respect to !and get d d![1 (!2 n !2)2+
2!2] =2![2(!2 n !2)
2] [(!2 n !2)2+
2!2]2(B.1) 17Dierentiating equation (11) with respect to !we get ndn d! d d!= !!2 p[(!2 n !2)2+
2!2] +!!2 p(!2 n !2)[2(!2 n !2)
2] [(!2 n !2)2+
2!2]2 =!!2 p[(!2 n !2)2 !2 n
2] [(!2 n !2)2+
2!2]2(B.2) Similarly dierentiating equation (12) with respect to !we get 2dn d!+ 2nd d!=!2 p
[(!2 n !2)2+
2!2] + 2!2!2 p
[2(!2 n !2)
2] [(!2 n !2)2+
2!2]2 =!2 p
(!2 n !2)[(!2 n !2) + 4!2] !2!2 p
3 [(!2 n !2)2+
2!2]2:(B.3) 18REFERENCES 1. Chakrabarti S 2006 Phys. Educ. 23167- 175 (New Delhi, India: South Asian Publishers PVT LTD) 2. Christy R W 1972 Am.J.Phys. 401403-1419 3. Kitamura R, Pilon L and Jonasz M 2007 Applied Optics 46(33) 8118-8133 4. Almog I F, Bradley M S and Bulovic V 2011 The Lorentz Oscillator and its Applications ( MIT OpenCourseWare, MIT6-007S11/lorentz) 5. Born M, Wolf E 1980 Principles of Optics ( New York:Pergamon Press, Sixth.Edn.) p 85 6. Kittel C 1976 Introduction To Solid State Physics (New Delhi:WileyEastern Limited, 5th Edn.) p 405 7. Hecht E 2002 Optics ( Delhi,India:Pearson Education, 4th Edn.) p 71, 85,70,128 8. Feynman R P, Leighton R B, Sands M 2003 The Feynman Lectures on Physics, 2nd Volume (New Delhi,India:Narosa Publishing House) p 1211,1210 9. Oughstun K E, Cartwright N A 2003 Optics Express 11(13) 1541-1546 10. Tanner D B 2013 Optical eects in solids (www.phys.u
.edu/tanner/notes.pdf ) 11. Kleppner D, Kolenkow R J 1973 An Introduction To Mechanics (New Delhi :Tata Mcgraw Hill ) p 426 12. Heitler W 1954 The quantum theory of radiation (Oxford University Press, 3rd. Edn.) p 163 1913. Seitz F 1940 The Modern Theory of Solids (McGraw-Hill ,International Series In Pure And Applied Physics) p 629 14. Jackson J D 1999 Classical Electrodynamics 3rd edn. (John Wiley & Sons, Inc) p 314 15. Jenkins F A , White H E 1957 Fundamentals of Optics (Mcgraw-Hill book company,Inc, 3rd Edn.) p 486 16. Sigel G H Jr 1973/74 Journal of Non-Crystalline Solids 13372-398 20Table 1: Refractive indices as a function of wavelengths for the
int glass prism [1] wavelength refractive index (nm) n 706.544 1.6087 667.815 1.6108 587.574 1.6167 504.774 1.6259 501.567 1.6264 492.193 1.6277 471.314 1.6311 447.148 1.6358 438.793 1.6377 21Table 2: Parameters obtained from tting of data of refractive indices to Lorentz model[1] !n N ! p !0 rad/s m 3rad/s rad/s 14:510151:02102918:0101517:81015 Table 3: Table for
t,!tand! f=
t !0
t !t ! rad/s rad/s rad/s 0.65015 11 :6101516:8101515:11015 Table 4: Parameters at the true resonance frequency frequency n 1 g(!) =4mE(!) (qE0 0)2 ! m 1m s2=rad2 !t 0.995 0.729 8 :181070.0122 0:15310 31 Table 5: Values of some parameters in the ultraviolet region frequency n 1 g(!) =4mE(!) (qE0 0)2 ! m 1m s2=rad2 !0 0.906 0.678 8 :061070.0124 0:14910 31 ! 1.18 0.769 7 :761070.0129 0:14110 31 !n 1.26 0.763 7 :391070.0135 0:13310 31 22 7.2 7.4 7.6 7.8 8 8.2 8.4 0.8 0.85 0.9 0.95 1 1.05α(107 m-1) ω/ω0Figure 1: Distribution of the absorption coecient in the absorption region 23 0.7 0.71 0.72 0.73 0.74 0.75 0.76 0.77 0.78 0.75 0.8 0.85 0.9 0.95 1κ ω/ω0Figure 2: Distribution of the extinction coecient in the absorption region 24 0.115 0.1155 0.116 0.1165 0.117 0.1175 0.118 0.1185 0.8 0.82 0.84 0.86 0.88 0.9 0.92 0.94 0.96 0.98 1 1.02R ω/ω0Figure 3: Distribution of Re
ectance Rin the absorption region 25 | 2019-07-10 | The Lorentz-Maxwell model of dispersion of light has been analyzed in this
paper to determine the true resonance frequency in the ultraviolet for the
electrons in optically transparent glasses and the damping coefficient at this
frequency. For this we needed the refractive indices of glass in the optical
frequency range. We argue that the true resonance condition in the absorption
region prevails when the frequency at which the absorption coefficient is
maximum is the same as the frequency at which the average energy per cycle of
the electrons is also a maximum. We have simultaneously solved the two
equations obtained from the two maxima conditions numerically to arrive at a
unique solution for the true resonance frequency and the damping coefficient at
this frequency. Assuming the damping coefficient to be constant over a small
frequency range in the absorption region, we have determined the frequencies at
which the extinction coefficient and the reflectance are maxima. These
frequencies match very well with the published data for silica glasses
available from the literature. | Determination of the damping co-efficient of electrons in optically transparent glasses at the true resonance frequency in the ultraviolet from an analysis of the Lorentz-Maxwell model of dispersion | 1907.04499v1 |
DRAFT VERSION JANUARY 24, 2022 Typeset using L ATEXpreprint2 style in AASTeX63 Cosmic ray streaming in the turbulent interstellar medium SIYAO XU1AND ALEX LAZARIAN2, 3 1Institute for Advanced Study, 1 Einstein Drive, Princeton, NJ 08540, USA; sxu@ias.edua 2Department of Astronomy, University of Wisconsin, 475 North Charter Street, Madison, WI 53706, USA; lazarian@astro.wisc.edu 3Centro de Investigaci ´on en Astronom ´ıa, Universidad Bernardo O’Higgins, Santiago, General Gana 1760, 8370993,Chile ABSTRACT We study the streaming instability of GeV 100GeV cosmic rays (CRs) and its damping in the turbulent interstellar medium (ISM). We find that the damping of streaming instability is dominated by ion-neutral col- lisional damping in weakly ionized molecular clouds, turbulent damping in the highly ionized warm medium, and nonlinear Landau damping in the Galactic halo. Only in the Galactic halo, is the streaming speed of CRs close to the Alfv ´en speed. Alfv ´enic turbulence plays an important role in both suppressing the streaming in- stability and regulating the diffusion of streaming CRs via magnetic field line tangling, with the effective mean free path of streaming CRs in the observer frame determined by the Alfv ´enic scale in super-Alfv ´enic turbu- lence. The resulting diffusion coefficient is sensitive to Alfv ´en Mach number, which has a large range of values in the multi-phase ISM. Super-Alfv ´enic turbulence contributes to additional confinement of streaming CRs, irrespective of the dominant damping mechanism. 1.INTRODUCTION The resonant streaming instability (Wentzel 1974; Kul- srud & Pearce 1969; Wentzel 1969; Skilling 1971) is im- portant for confining cosmic rays (CRs) with energies up to 100GeV in the Galaxy (Farmer & Goldreich 2004). It has many astrophysical implications on, e.g., shock accel- eration (Bell 1978), heating of intracluster media (Guo & Oh 2008; Brunetti & Jones 2014), launching galactic winds (Ipavich 1975; Wiener et al. 2017; Mao & Ostriker 2018; Holguin et al. 2019; Quataert et al. 2021), transport of CRs in starburst galaxies (Krumholz et al. 2020) and around CR sources (Marcowith et al. 2021), and explaining PAMELA and AMS-02 observations at Earth (Blasi et al. 2012; Amato & Casanova 2021). The self-generated Alfv ´en waves by CRs via the streaming instability are subject to various damping effects, including ion-neutral collisional damping in a partially ionized medium (Kulsrud & Pearce 1969; Plotnikov et al. 2021; Armillotta et al. 2021), nonlinear Landau damping in a collisionless medium (Kulsrud 2005), as well as turbulent damping by background Alfv ´enic turbulence (Lazarian 2016). Unlike other damping mechanisms depending on plasma conditions, turbulent damping depends on properties of magnetohydro- dynamic (MHD) turbulence. Measurements in different in- terstellar phases reveal a large range of turbulence parame- ters, e.g., Alfv ´en Mach number MAthat characterizes the magnetization level of turbulence (Lazarian et al. 2018; Hu et al. 2019). Based on the theoretical understanding of MHD turbulence developed since Goldreich & Sridhar (1995) and Lazarian & Vishniac (1999), Farmer & Goldreich (2004) first formulated aHubble Fellowthe turbulent damping rate for trans-Alfv ´enic (MA= 1) tur- bulence. Lazarian (2016) further provided a detailed analysis on turbulent damping in both super-Alfv ´enic (MA>1) and sub-Alfv ´enic (MA<1) turbulence. When the growth of streaming instability is limited by turbulent damping, the re- sulting streaming speed of CRs can deviate from the Alfv ´en speed and is sensitive to turbulence parameters. In addition, due to the magnetic field line tangling in super-Alfv ´enic tur- bulence, CRs streaming along turbulent magnetic fields have an effective mean free path determined by the Alfv ´enic scale lA=LM 3 A(Lazarian 2006; Brunetti & Lazarian 2007), whereLis the injection scale of turbulence, and an isotropic distribution on scales larger than lA. The above effect on the spatial diffusion of streaming CRs has not been addressed in previous studies. In this work, we focus on the effect of Alfv ´enic turbulence on the streaming speed and diffusive propagation of stream- ing CRs in the energy range GeV 100GeV in different tur- bulence regimes. We also examine the relative importance between turbulent damping and other damping mechanisms of streaming instability in various interstellar phases. In par- ticular, in a partially ionized medium, as MHD turbulence is also subject to ion-neutral collisional damping (Xu et al. 2015, 2016; Xu & Lazarian 2017a), the relative importance between turbulent damping and ion-neutral collisional damp- ing of CR-driven Alfv ´en waves depends on the ionization fraction and the coupling state between ions and neutrals in different ranges of length scales. The paper is organized as follows. The description on streaming instability and different damping effects is pre- sented in Section 2. In Section 3, we compare turbulent damping and ion-neutral collisional damping in both weakly and highly ionized media, and we derive the correspond- ing streaming speed and diffusion coefficient in different regimes. The comparison between turbulent damping andarXiv:2112.06941v2 [astro-ph.HE] 20 Jan 20222 nonlinear Landau damping in the Galactic halo is carried out in Section 4. Discussion and our summary are in Section 5 and Section 6, respectivley. 2.GROWTH AND DAMPING OF CR-DRIVEN ALFV ´EN WA VES 2.1. Growth of Alfv ´en waves The same resonance condition, rL, applies to both gyroresonant scattering of CRs by Alfv ´en waves and genera- tion of Alfv ´en waves via the CR resonant streaming instabil- ity, whereis the wavelength of Alfv ´en waves, and rLis the Larmor radius of CRs. For CRs streaming from a source to a sink, when their bulk drift velocity, i.e., streaming velocity vD, is larger than the Alfv ´en speedVA, the Alfv ´en waves ex- cited by streaming CRs become unstable. The wave growth rate is (Kulsrud & Pearce 1969) CR= 0nCR(>rL) nvD VA 1 ; (1) when neutrals and ions are strongly coupled together with the Alfv ´en wave frequency r 1 LVAmuch smaller than the neutral-ion collisional frequency ni=
diin a weakly ion- ized medium or the ion-neutral collisional frequency in=
dnin a highly ionized medium. Here
dis the drag coeffi- cient (Shu 1992), iandnare the ion and neutral mass den- sities, 0=eB0=(mc)is the nonrelativistic gyrofrequency, eandmare the proton electric charge and mass, cis the light speed,nCR(> rL)is the number density of CRs with the Larmor radius larger than rL,nis the total number den- sity of gas,vD VAis the drift velocity in the wave frame, VA=B0=p4,B0is the mean magnetic field strength, and=i+nis the total mass density. When neutrals and ions are weakly coupled with r 1 LVAi> inin a partially ionized medium, where VAi= B0=p4iis the Alfv ´en speed in ions, or in a fully ionized medium, the growth rate is CR= 0nCR(>rL) nivD VAi 1 : (2) Hereniis the ion number density. The CR-generated Alfv ´en waves in turn scatter the CRs. The quasilinear gyroresonant scattering of CRs in the wave frame regulates vD VA(i). In a steady state, the ampli- tude of CR-driven Alfv ´en waves is stabilized by the balance between CRand the damping rate of Alfv ´en waves. The pitch-angle scattering corresponding to this wave amplitude is also in balance with the net streaming (Kulsrud 2005). The net drift velocity in the wave frame in a steady state is (Kul- srud 2005; Wiener et al. 2013) vD VA(i)=1 3vrL HB2 0 B(rL)2; (3) wherevcfor relativistic CRs, His the distance from the source to the sink, and B(rL)2=B2 0is the relative magnetic fluctuation energy of the resonant Alfv ´en waves.The damping of streaming instability depends on both properties of the background MHD turbulence and plasma conditions of the surrounding medium. Next we will discuss different damping mechanisms. 2.2. Turbulent damping Turbulent damping was first mentioned in Yan & Lazar- ian (2002) and later studied in detail by Farmer & Goldre- ich (2004) for trans-Alfv ´enic turbulence and Lazarian (2016) in various turbulence regimes for a more general astrophysi- cal application. For strong MHD turbulence with the critical balance (Goldreich & Sridhar 1995) between the turbulent motion in the direction perpendicular to the local magnetic field and the wave-like motion along the local magnetic field (Lazarian & Vishniac 1999), i.e., x? ux=xk VA; (4) wherex?andxkare the length scales of a turbulent eddy perpendicular and parallel to the local magnetic field, and ux=Vst(x?=Lst)1 3 (5) is the turbulent velocity at x?. The corresponding turbulent cascading rate, i.e., eddy turnover rate, is uxx 1 ?=VstL 1 3 stx 2 3 ?: (6) Here Vst=VA; Lst=lA=LM 3 A; (7) for super-Alfv ´enic turbulence with the Alfv ´en Mach number MA=VL=VA>1,lAis the Alfv ´enic scale, and Vst=VLMA; Lst=ltran=LM2 A; (8) for sub-Alfv ´enic turbulence with MA<1, whereVLis the turbulent velocity at the injection scale Lof turbulence. We follow the analysis in Lazarian (2016) to derive the tur- bulent damping rate. The CR-driven Alfv ´en waves propagate along the local magnetic field. For the Alfv ´en waves with the wavelength , the distortion by the turbulent motion at the resonant perpendicular scale x?is most efficient. andx? are related by x? VA= ux: (9) The scaling relations in Eqs. (4) and (9) are illustrated in Fig. 1, and they give =ux VAx?=u2 x V2 Axk: (10) By inserting Eq. (5) into Eq. (9), one finds x?=3 4VA Vst3 4L1 4 st: (11)3 Figure 1. Sketch of the relation between xkandx?for strong anisotropic MHD turbulence and the relation between x?andfor turbulent damping of CR-driven Alfv ´en waves. The turbulent damping rate is determined by the eddy turnover rate at x?(Eqs. (6) and (11)), st=ux x?=V 1 2 AV3 2 stL 1 2 st 1 2: (12) Note thatx?should lie within the range of strong MHD tur- bulence, i.e., [xmin;?;Lst], wherexmin;?is the perpendicular damping scale of MHD turbulence and determined by micro- scopic plasma effects. The corresponding range of rLis (Eq. (11)), Vst VAL 1 3 stx4 3 min;?<rL<Vst VALst: (13) Eqs. (12) and (13) become (Eq. (7)) st=VAL 1 2M3 2 A 1 2=VLL 1 2M1 2 A 1 2; (14) and l 1 3 Ax4 3 min;?<rL<lA; (15) for super-Alfv ´enic turbulence, and (Eq. (8)) st=VAL 1 2M2 A 1 2=VLL 1 2MA 1 2; (16) and L 1 3M4 3 Ax4 3 min;?<rL<LM4 A; (17) for sub-Alfv ´enic turbulence. We see that stincreases with MA. Naturally, a larger amplitude of turbulence can result in a more efficient turbulent damping. For the same reason, stof sub-Alfv ´enic turbulence is smaller than that of super- Alfv ´enic turbulence under the same physical condition. 2.3. Ion-neutral collisional damping in a partially ionized medium Alfv ´en waves propagating in the partially ionized interstel- lar medium (ISM) with a wide range of ionization fractions, e.g., from weakly ionized molecular clouds (MCs) to highlyionized warm phases, are subject to the damping effect due to the collisional friction between ions and neutrals. In a weakly ionized medium with ni< in, when ions and neutrals are strongly coupled together with the wave fre- quency!=VAkk< ni, the ion-neutral collisional (IN) damping rate is (Piddington 1956; Kulsrud & Pearce 1969) IN=nV2 Ak2 k 2ni; (18) wherekkis the wavevector component parallel to the mag- netic field, and n=n=. When neutrals and ions are de- coupled from each other, i.e., in the weak coupling regime with!=VAikk>in, there is IN=in 2: (19) MHD turbulent cascade in a weakly ionized medium is also subject to IN damping (Xu et al. 2015, 2016; Xu & Lazarian 2017a). We consider that the driving of turbulence occurs in the strong coupling regime. MHD turbulence is damped when INin Eq. (18) equalizes with the turbulent cascading rate ukk?, whereukis the turbulent velocity at wavenumber k, andk?is the wavevector component perpen- dicular to the magnetic field. For strong MHD turbulence, k? andkkare related by the critical balance relation (see Section 2.2) k?uk=kkVA: (20) The corresponding IN damping scale of MHD turbulence is (Xu et al. 2015, 2016) xmin;?=2ni n 3 2L 1 2 stV3 2 st; (21) which gives the smallest perpendicular scale of MHD turbu- lent cascade. It becomes xmin;?=2ni n 3 2L 1 2V3 2 L (22) for super-Alfv ´enic turbulence, and xmin;?=2ni n 3 2L 1 2V3 2 LM1 2 A (23) for sub-Alfv ´enic turbulence. With ukk?=VAkk<ni<in; (24) and nV2 Ak2 k 2ni<nni 2<ni 2<in 2; (25) strong MHD turbulence injected in the strong coupling regime cannot cascade into the weak coupling regime, and INof Alfv ´en waves in the weak coupling regime is larger than INand the eddy turnover rate of MHD turbulence in the strong coupling regime (Xu et al. 2016).4 In a highly ionized medium with in< ni, in the strong coupling regime with VAkk<in, INis given by Eq. (18). When ions are decoupled from neutrals with VAikk> in, there is (Xu et al. 2016) IN=niV2 Aik2 k 2 (1 +)22 ni+V2 Aik2 k; (26) where=n=i. When neutrals and ions are decoupled from each other with VAikk>ni, the above expression can be reduced to Eq. (19). As ukk?=VAkk(orVAikk)> IN in both strong and weak coupling regimes, MHD turbulence in a highly ionized medium is not damped by IN damping. Briefly, IN damping is sensitive to the ionization fraction, and the damping effect in a weakly ionized medium is much stronger than that in a highly ionized medium. 2.4. Nonlinear Landau damping In the fully ionized gaseous Galactic halo or corona (Spitzer 1990; McKee 1993), Alfv ´en waves are subject to nonlinear Landau (NL) damping due to the resonant interac- tions of thermal ions with the beat waves produced by cou- ples of Alfven waves (Lee & V ¨olk 1973; Kulsrud 1978). The damping rate is (Kulsrud 1978) NL=1 2 21 2vth cB(rL)2 B2 0 ; (27) where =eB0=(
mc)c=rLis the gyrofrequency of rel- ativistic CRs with the Lorentz factor
,vth=p kBTi=mi is the average thermal ion speed, kBis the Boltzmann con- stant,Tiis ion temperature, and miis ion mass. Unlike st and IN, NLdepends on the amplitude of CR-generated Alfv ´en waves. 3.TURBULENT DAMPING VS. IN DAMPING Depending on the driving condition of MHD turbulence and the plasma condition in different interstellar phases, the dominant damping mechanism of streaming instability varies. We first compare turbulent damping with IN damp- ing in weakly and highly ionized media, and then compare turbulent damping with NL damping in a fully ionized hot medium (see Section 4). As the streaming instability and wave damping together determine vD, a proper description of the damping effect in different regimes is important for determining the diffusion coefficient of CRs and understand- ing their confinement in the Galaxy. 3.1. Dominant damping mechanism in different regimes (1) Weakly ionized medium. We first consider the case when both MHD turbulence and CR-driven Alfv ´en waves are in the strong coupling regime, i.e., r 1 LVA< ni. If the turbulent damping is the dominant damping mechanism, we should have (i): st(x?)> IN(xk); (28) so that MHD turbulence is not damped at x?, and (ii): st(x?)> IN(rL): (29)We easily see rL<x?<xk (30) based on the relation in Eq. (10), meaning IN(rL)> IN(xk): (31) Therefore, if condition (ii) is satisfied, then condition (i) is naturally satisfied. As an example, using the following parameters, we have VA rLni =0:07B0 1G2nH 100cm 3 3 2ne=nH 0:1 1ECR 10GeV 1 <1; (32) wherene=nHis the ionization fraction, neandnHare num- ber densities of electrons and atomic hydrogen, mi=mn= mH,mnis neutral mass, mHis hydrogen atomic mass,
d= 5:51014cm3g 1s 1(Shu 1992), and ECRis the energy of CR protons. The values used here do not represent the typical conditions of MCs, but are still considered as a possibility given the large variety of interstellar conditions. Condition (ii) in Eq. (29) can be rewritten as (Eqs. (14) and (18)) MA>n 2niVAL1 2r 3 2 L2 3 = 2B0 1G5 3nH 100cm 3 1ne=nH 0:1 2 3 L 0:1pc1 3ECR 10GeV 1(33) for super-Alfv ´enic turbulence driven on small length scales, e.g., near supernova shocks when the shock and shock pre- cursor interact with interstellar or circumstellar density inho- mogeneities (e.g., Xu & Lazarian 2017b, 2021). We note that the outer scale of this turbulence is determined by the size of density clumps. For instance, the typical size of ubiquitous HI clouds in the ISM is 0:1pc (Inoue et al. 2009). As this scale is much larger than rLof low-energy CRs considered here, the CR-induced Alfv ´en waves are subject to turbulent damping in this scenario. With the above parameters used, in Fig. 2(a), the shaded area shows the ranges of MAandne=nHfor turbulent damp- ing to dominate over IN damping. The solid line represents MA=n 2niVAL1 2r 3 2 L2 3; (34) below which, IN damping dominates over turbulent damping. In the area above the solid line, as MHD turbulence is also subject to IN damping, to ensure that the condition in Eq. (15) is also satisfied, other constraints on MAindicated in Fig. 2(a) are MA<2ni nL VA1 3; (35)5 corresponding to (Eq. (22)) xmin;?<lA; (36) MA<h2ni n2 V 2 ALrLi1 3; (37) corresponding to (Eqs. (11) and (22)) st(x?)> IN(xk); (38) and MA<L rL1 3; (39) corresponding to rL<lA: (40) In addition, the vertical dashed line indicates the ne=nH value corresponding to r 1 LVA=ni. Toward a larger ne=nH, the Alfv ´en waves are in the strong coupling regime. In Fig. 2(b), using the MAvalue given by Eq. (34), we illustrate the relation between different length scales. For the regime of interest, we have rL;min<VA ni<rL<x?<xk<lA; (41) whererL;min=l 1 3 Ax4 3 min;?is given in Eq. (15). In typical MC conditions, we find that CR-driven Alfv ´en waves are in the weak coupling regime with VAi rLin2103B0 10G2nH 100cm 3 3 2ne=nH 10 4 1 2 ECR 1GeV 1 1: (42) For MHD turbulence injected at a large scale in the strong coupling regime, there is always (see Section 2.3) st(x?)< |