Abstract
Topology is quickly becoming a cornerstone in our understanding of electronic systems. Like their electronic counterparts, bosonic systems can exhibit a topological band structure, but in real materials it is difficult to ascertain their topological nature, as their ground state is a simple condensate or the vacuum, and one has to rely instead on excited states, for example a characteristic thermal Hall response. Here we propose driving a topological magnon insulator with an electromagnetic field and show that this causes edge mode instabilities and a large nonequilibrium steadystate magnon edge current. Building on this, we discuss several experimental signatures that unambiguously establish the presence of topological magnon edge modes. Furthermore, our amplification mechanism can be employed to power a topological travellingwave magnon amplifier and topological magnon laser, with applications in magnon spintronics. This work thus represents a step toward functional topological magnetic materials.
Similar content being viewed by others
Introduction
While fermionic topological insulators have a number of clear experimental signatures accessible through linear transport measurements^{1,2}, noninteracting bosonic systems with topological band structure have a simple condensate or the vacuum as their ground state^{3}, making it more difficult to ascertain their topological nature. Their excited states, however, may carry signatures of the topology of the band structure, for example, in form of a thermal Hall response^{4,5,6}. There is great interest in certifying and exploiting topological edge modes in bosonic systems, as they are chiral and robust against disorder, making them a great resource to realize backscatteringfree waveguides^{7,8} and potentially topologically protected travellingwave amplifiers^{9}. It has been predicted that topological magnon insulators (TMI) are realized, e.g., in kagome planes of certain pyrochlore magnetic insulators as a result of DzyaloshinskiiMoriya (DM) interaction^{4,10,11}. To date, there exists only indirect experimental proof, via neutron scattering measurements of the bulk band structure in Cu[1,3benzenedicarboxylate (bdc)]^{12}, and observation of a thermal magnon Hall effect in Lu_{2}V_{2}O_{7}^{5} and Cu(1,3bdc)^{6}. The main obstacle is that magnons are uncharged excitations and thus invisible to experimental tools like STM or ARPES with spatial resolution. An unambiguous experimental signature, such as the direct observation of an edge mode in the bulk gap is hampered by limitations in energy resolution (in resonant Xray scattering) or signal strength (in neutron scattering)^{12}.
Here, we propose driving a magnon edge mode to a parametric instability, which, when taking into account nonlinear damping, induces a nonequilibrium steady state with a large chiral edge mode population. In such a state the local polarization and magnetization associated with the edge mode are coherently enhanced, which could enable direct detection of edge modes via neutron scattering. Crucially, we show that selective amplification of edge modes can be achieved while preserving the stability of the bulk modes and thus the magnetic order. Another key experimental signature we predict is that applying a driving field gradient gives rise to a temperature gradient along the transverse direction, thus establishing what one might call a driven Hall effect (DHE). Topological magnon amplification has further uses in magnon spintronics^{13}, providing a way to amplify magnons and to build a topological magnon laser^{14,15}. Our work on driving topological edge modes in magnetic materials complements previous investigations in ultracold gases^{16,17}, photonic crystals^{9}, and most recently arrays of semiconductor microresonators^{14,15} and graphene^{18}.
Results
Edge mode parametric instability
Before discussing a microscopic model, we show qualitatively how a parametric instability may arise from anomalous magnon pairing terms in a chiral onedimensional waveguide. We consider bosonic modes \(\{ \hat a_k\}\) with energies {ω_{k}} (for example the magnon edge mode between the first and second band, as in Fig. 1), labelled by momentum k, interacting with another bosonic mode (electromagnetic field mode) \(\hat b\), as described by generic threewave mixing (ħ = 1)
Under strong coherent driving, the bosonic annihilation operator \(\hat b\) can be replaced by its classical amplitude \(\hat b \approx \beta \,{\mathrm{exp}}(  i{\mathrm{\Omega }}_0t) \gg 1\), yielding an effective Hamiltonian \(\hat H = \hat H_0 + g_k\beta [\hat a_{  k}^\dagger \hat a_k^\dagger {\mathrm{exp}}(i{\mathrm{\Omega }}_0t) + \hat a_k\hat a_{  k}{\mathrm{exp}}(  i{\mathrm{\Omega }}_0t)]\). The second term produces magnon pairs with equal and opposite momentum. The timedependence can be removed by passing to a rotating frame with respect to \(\mathop {\sum}\nolimits_k {({\mathrm{\Omega }}_0/2)} \hat a^\dagger_k\hat a_k\). From the Hamiltonian, it is straightforward to derive the equations of motion, which couple particles at momentum k with holes at −k. Neglecting fluctuations, we focus on the classical amplitudes of the fields \(\alpha _k = \langle \hat a_k\rangle\) and include a phenomenological linear damping rate γ_{k} to take into account the various damping processes present in such materials^{12,19}. As we are interested in amplification around a small bandwidth, we neglect the momentum dependence of the coupling g_{k} ≃ g and damping γ_{k} ≃ γ, arriving at
where we have introduced the frequency relative to the rotating frame \(\tilde \omega _k = \omega _k  {\mathrm{\Omega }}_0/2\), the vector \({\mathbf{A}}_k = (\alpha _k,\alpha _{  k}^ \ast )\), and the overall coupling strength \({\cal{E}} = g\beta\). The eigenvalues of the dynamical matrix Eq. (2) are the complex energies
If the coupling \({\cal{E}}\) exceeds the energy difference between pump photons and magnon pair (the detuning) \(\tilde \omega _k + \tilde \omega _{  k} = \omega _k + \omega _{  k}  {\mathrm{\Omega }}_0\), the square root becomes imaginary. If further its magnitude exceeds γ, more particles are created than dissipated, causing an instability and exponential growth of the number of particles in this mode. Eventually the growth is limited by nonlinear effects, as discussed below. Despite its simplicity, Eq. (3) provides an accurate account of the fundamental instability mechanism in twodimensional kagome TMIs, as is illustrated by the quantitative agreement Fig. 2. This forms the key ingredient for directly observing chiral magnon edge modes.
Microscopic model
Turning to a more realistic model, we consider spins on the vertices of an insulating kagome lattice ferromagnet that interact via Heisenberg and DzyaloshinskiiMoriya (DM) interaction
Here, D_{ij} is the DM vector that can in principle differ from bond to bond, but is heavily constrained by lattice symmetries. H_{0} is an externally applied magnetic field, μ_{B} the Bohr magneton, g_{L} the Landé gfactor, and J is the Heisenberg interaction strength. This model has been found to describe the thermal magnon Hall effect in Lu_{2}V_{2}O_{7}^{5}, as well as the bulk magnon band structure of Cu(1,3bdc)^{12}.
The low energy excitations around the ferromagnetic order are magnons, whose bilinear Hamiltonian is obtained from a standard HolsteinPrimakoff transformation to order 1/S along the direction of magnetization, i.e., \(\hat S^ + = \sqrt {2s} \hat a,\hat S^  = \sqrt {2s} \hat a^\dagger ,\hat S^z = s  \hat a^\dagger \hat a\)^{4,10}, yielding
where K_{0} is a constant, the sum ranges over bonds directed counterclockwise in each triangle, and we have chosen the magnetic field to point along z, introducing \(h \equiv g_{\mathrm{L}}\mu _{\mathrm{B}}H_0^z\).
To second order, the Hamiltonian only contains the component of D_{ij} along z (D_{z}), which is the same for all bonds due to symmetry. We take the unit cell to be one upright triangle (red in Fig. 1), with sites ρ_{1} = (0, 0), \(\rho _2 = (  1,\sqrt 3 )/4\), ρ_{3} = (−1/2, 0). The unit cells form a triangular Bravais lattice generated by the lattice vectors δ_{1} = (1, 0), \(\delta _2 = (1,\sqrt 3 )/2\), \(\delta _3 = \delta _2  \delta _1 = (  1,\sqrt 3 )/2\). For nonzero D_{z}, the bands in this model are topological^{10,12} causing exponentially localized edge modes to appear within the band gaps.
The effect of an oscillating electric field on magnons in a TMI is characterized by the polarization operator, which can be expanded as a sum of singlespin terms, products of two spins, three spins, etc^{20}. Lattice symmetries restrict which terms may appear in the polarization tensor^{20}. In the pyrochlore lattice, the polarization due to single spins (linear Stark effect) vanishes, as each lattice site is a centre of inversion, such that the leading term contains two spin operators. The associated tensor can be decomposed into the isotropic (trace) part π, as well as the anisotropic traceless symmetric and antisymmetric parts Γ and D, viz. \(\widehat {\mathbf{P}}_{jl} = (\pi _{jl}\delta ^{\beta \gamma } + {\mathbf{\Gamma }}_{jl}^{(\beta \gamma )} + {\mathbf{D}}_{jl}^{[\beta \gamma ]})\hat S_j^\beta \hat S_l^\gamma\) (sum over β, γ implied). Kagome TMIs generically have a nonzero anisotropic symmetric part, which implies the presence of anomalous magnon pairing terms in the spinwave picture
The polarization enters the Hamiltonian via coupling to the amplitude of the electric field, \(\hat H(t) = \hat H_0  {\mathbf{E}}(t) \cdot \widehat {\mathbf{P}}\), thus introducing terms that create a pair of magnons while absorbing a photon. Pair production of magnons is a generic feature of antiferromagnets (via π)^{20}, but since in ferromagnets it relies on anisotropy, it is expected to be considerably weaker. A microscopic calculation based on a thirdorder hopping process in the FermiHubbard model at half filling reveals that Q = ae(t/U)^{3}/2, where a is the lattice vector, e the elementary charge, t the hopping amplitude, and U the onsite repulsion (see Supplementary Note 1).
As in the chiral waveguide model, assume an oscillating electric field E(t) = E_{0} cos (Ω_{0}t). We consider an infinite strip with W unit cells along y, but remove the lowest row of sites to obtain a manifestly inversionsymmetric model. Note that this is a choice out of convenience and that inversion symmetry is by no means a requirement for our scheme. Diagonalizing the undriven Hamiltonian \(\hat H_0\) (4), we label the eigenstates b_{k,s} by their momentum along x and an index s ∈ {1, 2,⋯, 3W − 2}. After performing the rotatingwave approximation, the full Hamiltonian reads
where we have introduced \(\tilde \omega _{k,s} = \omega _{k,s}  {\mathrm{\Omega }}_0/2\), and \(\widetilde {\mathbf{Q}}_{ss\prime }(k)\), which characterizes the strength of the anomalous coupling between two modes. It is obtained from Q_{mn} through Fourier transform and rotation into the energy eigenbasis (cf. Supplementary Note 3).
As in the onedimensional waveguide model, a pair of modes is rendered unstable if their detuning Δ_{k,ss′} = ω_{k,s} + ω_{−k,s′} − Ω_{0} is smaller than the anomalous coupling between them. The detuning Δ_{k,ss′} varies quickly as a function of k except at points where the slopes of ω_{k,s} and ω_{−k,s′} coincide to first order, which happens at k =0, π when s = s′. At those values of k, the energy matching condition is fulfilled for a broader range of wavevectors, which leads to a larger amplification bandwidth. However, the edge modes are only localized to the edge around k = π, such that driving around k = π is most efficient, which we consider here (cf. Fig. 1). Expanding the dispersion to second order around this point, we find \(\omega _{\pi + q} \simeq \omega _\pi + q\omega _ \pi ^ {\prime} + (q^2/2)\omega _ \pi ^ {\prime\prime} + {\cal{O}}(q^3)\), yielding \({\mathrm{\Delta }}_{\pi + q} = 2\omega _\pi  {\mathrm{\Omega }}_0 + q^2\omega _ \pi ^ {\prime\prime} + {\cal{O}}(q^4)\). Placing the pump at Ω_{0} = 2ω_{π} thus makes magnon pairs around k = π resonant, on a bandwidth of order \(\sqrt {{\cal{E}}/\omega {\prime\prime}_\pi}\). For weak driving, where the bandwidth is low, higherorder terms in the dispersion relation can be neglected, and this simple calculation captures the amplification behaviour extremely well, as we illustrate in Fig. 2. We calculate the band structure and find the unstable modes numerically (see Methods), with \(\tilde \omega _{k,s}\) and \(\widetilde {\mathbf{Q}}_{ss\prime }\) obtained from a microscopic model detailed in Supplementary Note 2, and plot the resulting band structure with instabilities in Fig. 1.
As we have seen above, an instability requires the anomalous terms to overcome the linear damping and the effective detuning. Linear damping, which we include as a uniform phenomenological parameter γ, has important consequences, as it sets a lower bound for the amplitude of the electrical field required to drive the system to an instability. It also ensures bulk stability. We have seen that there are three conditions for a parametric instability. First, there has to exist a pair of modes whose lattice momenta add to 0 (or 2π); Second, the sum of their energy has to match the pump frequency; and Third, the strength of their anomalous interaction has to overcome both their detuning and their damping. While momentum and energy matching is by design fulfilled by the edge mode, there is a large number of bulk mode pairs that also fulfil it. We show in Supplementary Note 4 that choosing the polarization to lie along γ increases the anomalous coupling for modes with wavevector close to π and that the coupling is small for almost all bulk mode pairs. The reason for this is that the bulk modes are approximately standing waves along y, and most bulk mode pairs have differing numbers of nodes, such that their overlap averages to zero. The remaining modes with appreciable anomalous coupling are far detuned in energy. This way, robust edge state instability can be achieved without any bulk instabilities, as demonstrated in Fig. 1. Bulk stability is crucial for the validity of the following discussion.
In the presence of an instability, the linear theory predicts exponential growth of edge magnon population. In a real system, the exponential growth is limited by nonlinear damping, for which we introduce another uniform parameter η, in the same spirit as Gilbert damping in nonlinear LandauLifshitzGilbert equations^{21}, such that Eq. (2) becomes
Microscopically, such damping arises from the next order in the spinwave expansion that allows fourwave mixing. While the linear theory only predicts the instability, Eq. (8) predicts a steadystate magnon occupation given through \(\alpha _{\pi + q}^2 = \eta ^{  1}(\sqrt {4{\cal{E}}^2  q^4(\omega {\prime\prime} _\pi )^2}  \gamma )\) (cf. Methods), which we show in Fig. 2b.
Experimental signatures
TMIs exhibit a magnonic thermal Hall effect at low temperatures^{4,5}. A similar effect occurs when the magnon population is not thermal, but a consequence of coherent driving, realizing a driven Hall effect (DHE).
We can calculate the steadystate edge magnon current from the occupation calculated above,
where \(\nu _{\pi + q} \simeq \omega _\pi ^ \prime + q\omega _ \pi ^ { \prime\prime }\) is the group velocity and \({\mathrm{\Lambda }} = \root {4} \of {{(4{\cal{E}}^2  \gamma ^2)/(\omega _ \pi ^ { \prime\prime} )^2}}\) is the range over which the steadystate population is finite (which coincides with the range over which the modes become unstable). While the integral can be done exactly (cf. Methods) an approximation within ±5% is given through
For \(2{\cal{E}} \gg \gamma\), a characteristic scaling of steadystate current with driving strength appears, \(J_{{\mathrm{SS}}} \propto {\cal{E}}^{3/2}\), distinct from the linear dependence one would expect for standard heating.
In Fig. 3b, c, we demonstrate that the steadystate edge current depends on the drive strength in a fashion that is well described by Eq. (10). The orderofmagnitude equilibration time can be estimated from the solution to \(\dot \alpha = (1/2)({\cal{E}}  \eta \alpha ^2)\alpha\), and for \(\eta /{\cal{E}} \gg 1\) it evaluates to \(t_{{\mathrm{eq}}}\sim {\cal{E}}^{  1}{\mathrm{log}}({\cal{E}}/\eta )\sim 10^4J^{  1}\) for our chosen values of \({\cal{E}}\) and η.
A DHE arises when a rectangular slab of size L_{x} × L_{y} is driven by a field with a gradient along y, as sketched in Fig. 3a. If L_{x}, L_{y} ≫ ν_{π}t_{eq}, the edges equilibrate to a steadystate magnon population governed by Eq. (10). The difference between the steadystate magnon currents on top and bottom edge corresponds to a net energy current \(J_{{\mathrm{net}}}^x\) along x, which to first order in the drive strength difference \({\mathrm{\Delta }}_y{\cal{E}}\) can be written^{22}
where one should note that in this nonequilibrium setting κ_{xy} is not a proper conductivity as in conventional linear response. The net edge current causes one side of the system to heat up faster, resulting in a temperature difference transverse to the gradient. As the edge magnons decay along the edge, the reverse heat current is carried by bulk modes. For small temperature differences the heat current follows the temperature gradient linearly and thus the averaged temperature difference \({\mathrm{\Delta }}_xT = {\int} {dy} [T(L_x,y)  T(0,y)]/L_y = J_{{\mathrm{net}}}^x/\kappa _{xx}\). The temperature difference can thus be written in terms of the applied field strength difference
As a word of caution, we note that this relation relies on several key assumptions. To begin with, temperature is in fact not well defined along the edge, as there is a nonequilibrium magnon occupation. Edge magnons decay at a certain rate into phonons, which can be modelled as heating of the phonon bath. If the equilibration time scale of the latter is fast compared to the heating rate through magnon decay, one can at least associate a local temperature to the phonons. Similarly, the bulk magnon modes can be viewed as a fast bath for the magnon edge mode and similar considerations apply. Even if these assumptions are justified, the two baths do not need to have the same temperature. Next, the heat conductivity associated to magnons and phonons differ in general, such that the κ_{xx} appearing in Eq. (12) can only be associated with the bulk heat conductivity if the temperatures of the two baths are equal. Some of these complications have been recognized to also play an important role in measurements of the magnon thermal Hall effect^{22}.
While the abovementioned concerns make quantitative predictions difficult, the DHE is easily distinguishable from the thermal Hall effect, due to the strong dependence of the temperature difference Δ_{x}T on drive frequency and polarization, as well as the fact that below the cutoff \(2{\cal{E}} = \gamma\) no instability occurs and that \(J_{{\mathrm{net}}} \propto \sqrt {{\cal{E}}_{{\mathrm{avg}}}}\) for \(2{\cal{E}}_{{\mathrm{avg}}} \gg \gamma\), rather than the linear dependence one would expect from standard heating. In certain materials such as Cu(1,3bdc), the appearance or disappearance of the topological edge modes can be tuned with an applied magnetic field^{12}, a property that could be used to further corroborate the results of such an experiment.
A number of other experimental probes might be used to certify a large edge magnon current and thus the presence of edge states. On the one hand, with a large coherent magnon population in a given mode, the local magnetic field and electric polarization associated to that mode will be enhanced. In particular, techniques that directly probe local magnetic or electric fields, such as neutron scattering^{12,23} or xray scattering, which to date are not powerful enough to resolve edge modes^{12}, would thus have a coherently enhanced signal, for example, by almost two orders of magnitude when taking the conservative parameters in Fig. 2. On the other hand, heterostructures provide a way to couple the magnons out of the edge mode into another material^{21}, for example one with a strong spin Hall effect, in which they can be detected more easily. In this setup, again the fact that the edge magnons have a large coherent population should make their signal easily distinguishable from thermal noise.
Material realizations
The model of a kagome lattice ferromagnet with DM interaction has been found to describe the thermal magnon Hall effect in Lu_{2}V_{2}O_{7}^{5}, as well as the bulk magnon band structure of Cu(1,3bdc)^{12,24}. These materials are in fact 3D pyrochlore lattices, which can be pictured as alternating kagome and triangular lattices along the [111] direction. However, their topological properties can be captured by considering only the kagome planes^{10,11,12} (shown in Fig. 1), thus neglecting the coupling between kagome and triangular planes. It has been suggested that the effect of the interaction may be subsumed into new effective interaction strengths^{11} or into an effective onsite potential^{10}. Typical values for strength of the DM and Heisenberg interactions lie between D/J ≈ 0.18^{12}, J ≈ 0.6 ± 0.1 meV ≃ 150 ± 30 GHz^{12} in Cu(1,3bdc) and D/J ≈ 0.32^{5}, J ≈ 3.4 meV ≈ 0.82 THz^{10} in Lu_{2}V_{2}O_{7}. The energy of the edge states close to k = π is approximately J, such that the applied drive needs to be at a frequency ω_{0}/2π = 0.3–1.6 THz. While experimentally challenging, low THz driving down to 0.6 THz has recently been achieved^{18,25}. Furthermore, the magnon energy can be tuned by applied magnetic fields.
An instability requires E_{0}ae(t/U)^{3} ≳ γ. With a ≃ 10Å^{10}, J ≃ 1 meV, and assuming t/U ≃ 0.1, we can estimate the minimum field strength required to overcome damping γ_{k} ≃ 10^{−4}J to be E_{0} ≃ 10^{5} V/m, although for quantitative estimates one would require both accurate values for the damping of the edge modes (at zero temperature) and t/U. This is accessible in pulsed operation^{18,25}, and perhaps in continuous operation through the assistance of a cavity.
Since the qualitative behaviour we describe can be derived from general and phenomenological considerations, we expect it to be robust and present in a range of systems, as long as they allow for anisotropy, i.e., if bonds are not centres of inversion. We thus expect that topological magnon amplification is also possible in recently discovered topological honeycomb ferromagnet CrI_{3}^{26}.
Discussion
We have shown that appropriate electromagnetic driving can render topological magnon edge modes unstable, while leaving the bulk modes stable. The resulting nonequilibrium steady state has a macroscopic edge magnon population. We present several strategies to certify the topological nature of the band structure, namely, implementing a driven Hall effect (DHE), direct detection with neutron scattering, or by coupling the magnons into a material with a spin Hall effect.
Our work paves the way for a number of future studies. As we have pointed out, edge mode damping plays an important role here. One might expect their damping to be smaller than that of generic bulk modes as due to their localization they have a smaller overlap to bulk modes. This suppression should be compounded by the effect of disorder^{21}, which may further enhance the feasibility of our proposed experiments. On the other hand, rough edges will have an influence over the matrix element between drive and edge modes, leading to variations in the anomalous coupling strength. Phonons in the material are crucial for robust thermal Hall measurements^{22} and could possibly mix with the chiral magnon mode^{27}, which motivates full microscopic calculations.
An exciting prospect is to use topological magnon amplification in magnon spintronics. There have already been theoretical efforts studying how magnons can be injected into topological edge modes with the inverse spin Hall effect^{21}. Given an efficient mechanism to couple magnons into and out of the edge modes, our amplification mechanism may enable chiral travellingwave magnon amplifiers, initially proposed in photonic crystals^{9}. Even when simply seeded by thermal or quantum fluctuations, the large coherent magnon steady state could power topological magnon lasers^{14}, with tremendous promise for future application in spintronics. In the near future, we hope that topological magnon amplification can be used for an unambiguous discovery of topological magnon edge modes.
Methods
Numerical calculation
For the numerical calculation, we choose a manifestly inversionsymmetric system obtained by deleting the lowest row of sites, a situation that is depicted in Fig. 1, where the tip of the lowest blue triangle is part of a unit cell whose other sites are not included. For example, repeating the star shown in Fig. 1 along x would result in an inversionsymmetric strip with W = 3. A Fourier transform of Eq. (5) along x yields a 3W − 2 by 3W − 2 Hamiltonian matrix for each momentum k
Note that we take ħ = 1. Diagonalizing this matrix yields singleparticle energy eigenstates with annihilation operator b_{k,s}, and a Hamiltonian \(H_0 = \mathop {\sum}\limits_{k,s} {\omega _{k,s}b^\dagger_{k,s}b_{k,s}}\). The resulting band structure is shown in Fig. 1. In our convention, the lowest bulk band has Chern number sgnD_{z}, the middle bulk band 0 and the top bulk band −sgnD_{z} (calculated, e.g., through the method described in ref. ^{28}). Accordingly, there is one pair of edge modes in each of the bulk gaps, one rightmoving localized at the lower edge and one leftmoving at the upper.
Including the anomalous terms obtained from a calculation based on the FermiHubbard model at half filling yields the full Hamiltonian Eq. (7). By means of a Bogoliubov transformation we obtain the magnon band structure and the unstable states^{29}, which form the basis for Fig. 2. The inclusion of nonlinear damping yields Eq. (8), which has been used to calculate Fig. 3. In the end, we calculate the current by evaluating the expectation value of the particle current or energy current operator, which can be obtained for a given bond from the continuity equation^{30}. The current across a certain cut of the system is obtained by summing the current operators for all the bonds that cross it. As the system we study is inversion symmetric, the total current in the x direction vanishes. In order to specifically find the edge current, we thus define a cut through half of the system, for example from the top edge to the middle.
Unstable modes in Bogoliubovde Gennes equation
We consider the full Hamiltonian
H_{0} − H_{rot} gives rise to the band structure shown in Fig. 1 above, while H_{amp} contains the anomalous terms. The idea of this section is to calculate which modes in Eq. (14) are unstable. Ideally, those should be the relevant edge modes, and only those. It turns out that this is possible in presence of linear damping.
Following ref. ^{9}, we define the vector \(a_k = (a_{k,1},a_{k,2}, \ldots ,a_{k,N},a^\dagger_{k_0  k,1}, \ldots ,a^\dagger_{k_0  k,N})^T\), where the index combines the label l_{y} and the site label in the unit cell and therefore runs from 1 to N = 3W − 2. The Hamiltonian can generically be written
where μ_{k} originates from H_{0} − H_{rot} and ν_{k} from H_{amp}. This form makes it evident that \(\mu _k = \mu _k^\dagger\) and \(\nu _k = \nu _k^T\). The equation of motion for this vector can be found from the Hamiltonian above and is
with σ_{z} = diag(1, 1, …, −1, −1, …) (N “ + 1”s and N “−1”s), with
We can then solve the eigenvalue problem and find stable and unstable modes. Furthermore, we can find the timeevolution for operators in the Heisenberg picture from Eq. (16). It is simply \(\left {a_k(t)} \right\rangle = e^{  i\sigma _zh_kt}\left {a_k(0)} \right\rangle\).
Steady state of nonlinear equations of motion
We start from the equations of motion (8) given in the main text, repeated here for convenience
In the steady state, \(\alpha _k^2 = {\mathrm{const}}.\), so we use the ansatz \(\alpha _k = {\mathrm{exp}}(i{\mathrm{\Delta }}t)\bar \alpha\), and \(\alpha _{  k}^ \ast = {\mathrm{exp}}(i{\mathrm{\Delta }}t)z\bar \alpha\) for some complex numbers z, \(\bar \alpha\) and real frequency Δ. As we are only interested in a narrow range of momenta, we expand the dispersion relation to second order, as in the main text. The pump frequency is set to match the edge mode at k = π, i.e., Ω_{0} = 2ω_{π}. As a consequence, \(\tilde \omega _{\pi + q} = q\omega _\pi ^\prime + q^2\omega _ \pi ^ {\prime\prime} /2 + {\cal{O}}(q^3)\).
If there is an instability, the solution \(\bar \alpha = 0\) is unstable. Assuming \(\bar \alpha \, \ne \, 0\) (thus \(z \, \ne \, 0\)), and for \({\mathrm{\Delta }} = q\omega _\pi ^\prime\), we find the set of equations
Multiplying the second equation by z^{2}, and subtracting the complex conjugate of the resulting equation from the first equation, one can show that z^{2} = 1. With this condition Eqs. (19) and (20) coincide, such that we can solve them for the intensity
This equation has solutions if and only if \(4{\cal{E}}^2 \ge q^4(\omega _\pi ^{\prime\prime} )^2 + \gamma ^2\), which coincides with the condition for the instability. If this condition is fulfilled, we have
The steadystate edge magnon current
where F(k, m) is the elliptic integral of the first kind.
Particle current operator
The particle current operator is obtained from the continuity equation for the number of magnons. We have
where h_{n} are local Hamiltonians defined through
The second term in Eq. (24) can be interpreted as a sum of the particle currents from n to the neighbouring sites m.
Data availability
Data sharing not applicable to this article as no datasets were generated or analysed during the current study.
References
Hasan, M. Z. & Kane, C. L. Colloquium: topological insulators. Rev. Mod. Phys. 82, 3045–3067 (2010).
Qi, X.L. & Zhang, S.C. Topological insulators and superconductors. Rev. Mod. Phys. 83, 1057–1110 (2011).
Vishwanath, A. & Senthil, T. Physics of threedimensional bosonic topological insulators: surfacedeconfined criticality and quantized magnetoelectric effect. Phys. Rev. X 3, 011016 (2013).
Katsura, H., Nagaosa, N. & Lee, P. A. Theory of the thermal Hall effect in quantum magnets. Phys. Rev. Lett. 104, 066403 (2010).
Onose, Y. et al. Observation of the magnon Hall effect. Science 329, 297–299 (2010).
Hirschberger, M., Chisnell, R., Lee, Y. S. & Ong, N. P. Thermal Hall effect of spin excitations in a kagome magnet. Phys. Rev. Lett. 115, 106603 (2015).
Haldane, F. D. M. & Raghu, S. Possible realization of directional optical waveguides in photonic crystals with broken timereversal symmetry. Phys. Rev. Lett. 100, 013904 (2008).
Wang, Z., Chong, Y., Joannopoulos, J. D. & Soljačić, M. Observation of unidirectional backscatteringimmune topological electromagnetic states. Nature 461, 772–775 (2009).
Peano, V., Houde, M., Marquardt, F. & Clerk, A. A. Topological quantum fluctuations and traveling wave amplifiers. Phys. Rev. X 6, 041026 (2016).
Zhang, L., Ren, J., Wang, J.S. & Li, B. Topological magnon insulator in insulating ferromagnet. Phys. Rev. B 87, 144101 (2013).
Mook, A., Henk, J. & Mertig, I. Magnon Hall effect and topology in kagome lattices: a theoretical investigation. Phys. Rev. B 89, 134409 (2014).
Chisnell, R. et al. Topological magnon bands in a kagome lattice ferromagnet. Phys. Rev. Lett. 115, 147201 (2015).
Chumak, A. V., Vasyuchka, V. I., Serga, A. A. & Hillebrands, B. Magnon spintronics. Nat. Phys. 11, 453–461 (2015).
Harari, G. et al. Topological insulator laser: Theory. Science 359, eaar4003 (2018).
Bandres, M. A. et al. Topological insulator laser: experiments. Science 359, eaar4005 (2018).
Galilo, B., Lee, D. K. & Barnett, R. Selective population of edge states in a 2D topological band system. Phys. Rev. Lett. 115, 245302 (2015).
Galilo, B., Lee, D. K. & Barnett, R. Topological edgestate manifestation of interacting 2D condensed bosonlattice systems in a harmonic trap. Phys. Rev. Lett. 119, 203204 (2017).
H. Plank, et al. Edge currents driven by terahertz radiation in graphene in quantum Hall regime, Preprint at http://arxiv.org/abs/1807.01525 (2018).
Chernyshev, A. L. & Maksimov, P. A. Damped topological magnons in the kagomelattice ferromagnets. Phys. Rev. Lett. 117, 187203 (2016).
Moriya, T. Theory of absorption and scattering of light by magnetic crystals. J. Appl. Phys. 39, 1042–1049 (1968).
Rückriegel, A., Brataas, A. & Duine, R. A. Bulk and edge spin transport in topological magnon insulators. Phys. Rev. B 97, 081106 (2018).
VinklerAviv, Y. & Rosch, A. Approximately quantized thermal Hall effect of chiral liquids coupled to phonons. Phys. Rev. X 8, 031032 (2018).
Yao, W. et al. Topological spin excitations in a threedimensional antiferromagnet. Nat. Phys. 14, 1011–1015 (2018).
Nytko, E. A., Helton, J. S., Müller, P. & Nocera, D. G. A structurally perfect S = 1/2 MetalOrganic Hybrid Kagomé Antiferromagnet. J. Am. Chem. Soc. 130, 2922–2923 (2008).
Karch, J. et al. Terahertz radiation driven chiral edge currents in graphene. Phys. Rev. Lett. 107, 276601 (2011).
Chen, L. et al. Topological spin excitations in honeycomb ferromagnet CrI_{3}. Phys. Rev. X 8, 041028 (2018).
Thingstad, E., Kamra, A., Brataas, A. & Sudbø, A. Chiral phonon transport induced by topological magnons. Phys. Rev. Lett. 122, 107201 (2019).
Fukui, T., Hatsugai, Y. & Suzuki, H. Chern numbers in discretized brillouin zone: efficient method of computing (Spin) Hall conductances. J. Phys. Soc. Jpn. 74, 1674–1677 (2005).
Blaizot, J. P. & Ripka, G. Quantum Theory of Finite Systems. (The MIT Press, Cambridge, MA, 1986).
Bernevig, B. A. & Hughes, T. L. Topological Insulators and Topological Superconductors, edition. 1sted (Princeton University Press, Princeton, New Jersey, 2013).
Acknowledgements
We are grateful to Ryan Barnett, Derek Lee, Rubén Otxoa, Pierre Roy, and Koji Usami for insightful discussions and helpful comments. D.M. acknowledges support by the Horizon 2020 ERC Advanced Grant QUENOCOBA (grant agreement 742102). AN holds a University Research Fellowship from the Royal Society and acknowledges support from the Winton Programme for the Physics of Sustainability and the European Union’s Horizon 2020 research and innovation programme under grant agreement No 732894 (FET Proactive HOT).
Author information
Authors and Affiliations
Contributions
D.M. performed the analysis and wrote the manuscript with assistance from J.K. and A.N. All authors contributed to the conception.
Corresponding author
Ethics declarations
Competing interests
The authors declare no competing interests.
Additional information
Peer review information: Nature Communications thanks Tomoki Ozawa and other anonymous reviewer(s) for their contribution to the peer review of this work.
Publisher’s note: Springer Nature remains neutral with regard to jurisdictional claims in published maps and institutional affiliations.
Supplementary information
Rights and permissions
Open Access This article is licensed under a Creative Commons Attribution 4.0 International License, which permits use, sharing, adaptation, distribution and reproduction in any medium or format, as long as you give appropriate credit to the original author(s) and the source, provide a link to the Creative Commons license, and indicate if changes were made. The images or other third party material in this article are included in the article’s Creative Commons license, unless indicated otherwise in a credit line to the material. If material is not included in the article’s Creative Commons license and your intended use is not permitted by statutory regulation or exceeds the permitted use, you will need to obtain permission directly from the copyright holder. To view a copy of this license, visit http://creativecommons.org/licenses/by/4.0/.
About this article
Cite this article
Malz, D., Knolle, J. & Nunnenkamp, A. Topological magnon amplification. Nat Commun 10, 3937 (2019). https://doi.org/10.1038/s41467019119142
Received:
Accepted:
Published:
DOI: https://doi.org/10.1038/s41467019119142
This article is cited by

NonHermitian dynamics and nonreciprocity of optically coupled nanoparticles
Nature Physics (2024)