Valley Chern Numbers and Boundary Modes in Gapped Bilayer Graphene
Electronic states at domain walls in bilayer graphene are studied by analyzing their four and two band continuum models, by performing numerical calculations on the lattice, and by using quantum geometric arguments. The continuum theories explain the distinct electronic properties of boundary modes localized near domain walls formed by interlayer electric field reversal, by interlayer stacking reversal, and by simultaneous reversal of both quantities. Boundary mode properties are related to topological transitions and gap closures which occur in the bulk Hamiltonian parameter space. The important role played by intervalley coupling effects not directly captured by the continuum model is addressed using lattice calculations for specific domain wall structures.
pacs:73.22.Pr, 77.55.Px, 73.20.-r
The electronic properties of few layer graphene systems depend sensitively on the atomic registry between neighboring layers MM (). For bilayer graphene (BLG) with stacking interlayer hybridization of orbitals on eclipsed lattice sites gaps out half of the low energy degrees of freedom, replacing the pseudorelativistic description of single layer graphene by a low energy theory in which quadratically dispersing chiral bands touch at discrete points in momentum space McCann (). A perpendicular electric field further breaks inversion symmetry and creates a semiconductor in which the gap size is determined by the electric field magnitude McCann (); Ohta (); Castro (); ZhangABC () and saturated at the strength of interlayer hybridization ZhangABC (). The possibility of exploiting this type of field tunable gap is being vigorously pursued in ultra-clean dual gated devices Yacoby1 (); Yacoby2 (); Henriksen (); Velasco (); Baoetal ().
More recently it has been appreciated that the field-induced gap admits a topological interpretation Martinetal (); SQH (). The low energy theory for BLG can be represented by an effective two-band model from which it is readily seen that inversion-symmetry breaking induces large momentum-space Berry curvatures SQH (); BerryRMP (). The Berry curvatures have opposite sign near the two inequivalent Brillouin-zone corners (valleys) at which gap is opened, so the integral of the Berry curvature over the full Brillouin zone is zero. Nonetheless, the integral of the Berry curvature within a single valley is nonzero and this allows a topological analysis of the valley-projected electronic spectrum. This idea has been developed in a continuum analysis of the subgap electronic states bound to a BLG domain formed by a sign reversal of the electric field between the layers Martinetal (); LiMorpurgo (); NatPhys (); Qiao (); Jung (). These electric-field walls (EFW) are predicted to bind pairs of subgap chiral co-propagating boundary modes, an interesting feature that can be related to the change in sign across a domain wall of a valley-projected topological index.
In this paper we examine a related BLG domain wall problem in which the interlayer electric field is uniform but the layer stacking switches from to . This new version of the problem changes the boundary conditions for matching the electronic states of the two bounding phases and requires that we augment the two-band model of BLG McCann (); Martinetal (); SQH () by accounting for all four sublattice degrees of freedom. Nonetheless we find that layer-stacking walls (LSW) bind electronic states with the same chiral structure as for the EFW studied previously. We make this connection explicit by mapping the two problems onto each other within a family of four-band BLG Hamiltonians. Our results demonstrate that the topological transition in a LSW structure is associated with a finite momentum gap closure in the parameter space of four-band BLG Hamiltonians. We construct a phase diagram (Fig. 1) which identifies the different types of topologically protected states that appear in BLG samples in which both the interlayer electric field and the layer stacking order vary in space. This analysis identifies yet a third type of domain wall in which the two pairs of chiral modes within a single valley are coupled, gapping the spectrum and annihilating boundary modes. Our results are supported by a continuum analysis of the domain wall states, lattice calculations for specific defect structures, and analysis using quantum geometrical arguments. Taken together these elements provide a general framework for understanding the origin of the valley-projected topological states in BLG, and their fragility in the presence of intervalley scattering.
The electronic states for BLG can be represented by four-component wavefunctions where denotes the atomic orbital centered on the or sites of the top or bottom layer. At low energies the Hamiltonian can be expanded for small around the two inequivalent Brillouin zone corners: with denoting and . Using Pauli matrices to represent operators that act on the sublattice degree of freedom within a layer and to represent operators acting on the layer degree of freedom, the BLG Hamiltonian can separated into layer diagonal and layer off-diagonal contributions by writing . We find that for stacked BLG in which the sites of top layer hybridize with the sites of the bottom layer, and where energies are in units of , , is the near-neighbor interlayer hopping amplitude, and is the electron velocity in an isolated layer. For the reversed stacking order the interlayer coupling term becomes . When present an electric potential difference adds with to the Hamiltonian.
When it is convenient to eliminate the high energy degrees of freedom at to arrive at an effective low energy two-band model McCann ()
where the matrices act on two component spinors in the low energy subspace for BLG and . Eq. (1) admits a geometrical interpretation in which the negative energy eigenstates are spinors aligned with and the filled band has a momentum space Berry curvature SQH (); BerryRMP ()
Because of the dependence in Eq. (2) the integral of over the full Brillouin zone is zero and the filled valence band carries total Chern number as required by time reversal symmetry. However, for small the Berry curvature is strongly peaked at the gap minima near and . Consequently, the integral of over an individual valley is accurately defined and the valley Chern number . The valley Chern number changes by across an EFW which can be associated with the appearance of pairs of valley-projected edge modes co-propagating along the boundary. These chiral modes have been obtained by analytic solution of the low energy two-band model in the presence of a sharp EFW and by numerical solution for a spatially varying that smoothly connects two electric-field reversed states Martinetal (). As noted in previous work SQH (); LiMorpurgo (), the introduction of a valley Chern number in this context is approximate since strictly speaking the construction does not map the full periodic Brillouin zone onto the parameter space of . Nonetheless, when is small and intervalley scattering is absent the computed change can be interpreted as a topological quantity, since is integrated over a closed surface produced by “gluing together” two integrals for the individual along a common boundary.
We now turn to the case of a LSW at which the bilayer registry reverses from local to local with held constant. Crossing a LSW changes the interlayer coupling matrix and switches the orbitals that span its low energy subspace. In this case evaluation of the Berry curvature requires consideration of all four degrees of freedom in the bilayer Dirac problem. Alternatively, one can identity the topological origin of LSW modes by examining the residual phase twists induced at large momentum in the eigenstates of the generalized Hamiltonian,
which reduces to the forms when . For degenerate single layer states deep in the filled band with energies are split by and are mixed by in the projected Hamiltonian
where are Pauli matrices acting in the subspace and . For large the eigenstates written in the original four orbital basis are
where and we explicitly display the overall phases . Using Eq. (5) we calculate the momentum space Berry connection
The change in the valley Chern numbers upon passing from the to states is obtained from the loop integral of the trace of over and band indices , which reads
Here are integer valued winding numbers of the overall phases . The dependence of this result vanishes after tracing over the filled bands, demonstrating that the valley Chern number in BLG is shared among all the occupied bands rather than being confined just to its low energy states as is often assumed. is a topological index provided that the difference is evaluated in the same gauge for the two bounding phases, which requires that for this boundary. It follows that and therefore that a domain wall separating insulating regions with local and registry will also confine pairs of valley-projected chiral modes propagating along the boundary with opposite velocities in the two valleys. Fig. 2 confirms this result by showing the spectra calculated by matching the full four component wavefunctions of Eq. (3) across a sharp boundary where switches from to .
Using Eq. (5) we find that at large the wavefunctions on the two sides of the LSW are related by a gauge transformation,
with a different phase twist induced in each layer. It follows from the accumulation of internal phases in the two layers that . EFW walls, where the sign of the potential difference between layers switches but the atomic registry does not change, can be analyzed similarly. Although does not appear explicitly in in Eq. (7) there is an implicit dependence through the phase prefactors. We find
i.e. that a sign reversal in can be absorbed in a sign change of combined with a change of basis. Using this construction the negative energy eigenstates at large on the two sides of the EFW are related by the following gauge transformation
which produces the same overall for the EFW. Eqs. (8)-(10) compactly express the relation between these two different types of domain wall in the four-band theory. This is also illustrated in Fig. 3 which compares the valley spectra computed for an EFW and a LSW showing their common chiral boundary modes.
In the LSW case, unlike the EFW case, analyzing the continuity of wavefunctions across the interface requires consideration of all four bands. The common topological origin of the domain wall spectra therefore becomes apparent only in a four-band continuum theory. Nevertheless, by integrating out the high energy bands at energies we are able to construct a two-band effective model away from the domain wall in which for either case is assigned to a sign change of the Berry curvature of the lower band. In this approach, Eq. (1) reads with layer Pauli matrices that act on different spinors in the cases: on for and on for . Because of inversion symmetry breaking, the valence band acquires a momentum space Berry curvature SQH (); BerryRMP (),
which integrates over a single valley to . Obviously, the valley Chern number changes by two across either a EFW or a LSW. Based on the bulk-boundary correspondence, pairs of valley-projected edge modes should co-propagate along the interface.
The plane phase diagram in Fig. 1 identifies distinct BLG topological phases. Phase boundaries occur along the and axes where the spectrum of undergoes gap closures at . describes an ungapped BLG system in which quadratic band crossing occurs exactly at , as seen in the left panel of Fig. 4. The gaps that open for the case of and are the electric field induced gaps easily understood within a two-band model. The boundary with and also has a gap closure, but it occurs at two finite momenta along the line where band crossing is possible because is a constant of the motion. For deviations in either or the degeneracies at the band-crossing points are lifted at linear order, implying the conical gap closure illustrated in the right panel of Fig. 4. When the original quadratic gap closure fissions into a pair of linear Dirac singularities each of which carries half the original winding number. Trajectories in the Hamiltonian parameter space that connect these topologically distinct ground states and involve different parameter values can shift the momenta at which the gap closures occur, but cannot eliminate them. For example, when but in Eq. (3) the layers decouple and the gap closure at degenerates to a closed Fermi ring with radius .
Fig. 1 also illustrates the possibility of a third type of compensated domain structure (labeled c) at which both the layer registry and interlayer electric field are reversed. Variation of local band parameters along this line connects two bilayer states that are distinct but have . As illustrated in the the lower panel of Fig. 3, spectra obtained by matching solutions across this compensated domain wall demonstrate that it hosts two pairs counter-propagating modes within the same valley, which hybridize and completely gap the spectrum.
The continuum model is able to explain the topological origin of the gapless interface modes. However, the short-range physics near the domain wall, which may be of essential significance, is not captured in the continuum Hamiltonian. Importantly, the single valley physics that protects the chiral domain wall solutions can be preempted by sufficiently strong large momentum scattering that acts to recouple states in the two valleys. In fact, Fig. 2 suggests that these single valley domain wall modes ultimately reconnect with each other. To study this further we construct a specific lattice model and use it to investigate how both lattice and interfacial effects,which couple the two valleys, influence the domain wall modes.
As depicted in Fig. 5, we consider the simplest LSW, i.e., a grain boundary separating BLG into left and right domains. Near the LSW, the lattices are continuous in one layer but fractured along a zigzag edge in the other. This introduces additional zigzag boundaries in the broken layer and allows switching of the bulk stacking order from () on the left to () on the right. For comparison, we first calculate the band structures for the case of uniform gapped BLG and for the case of gapped BLG with an EFW at which stacking order is preserved. As expected and shown in Fig. 6(a) and (b), quantum valley Hall edge states SQH () and two flat bands appear at the outer zigzag edges in uniformly gapped BLG. In the sample with an EFW there is an additional pair of co-propagating chiral gapless modes which emerge at each valley. Fig. 6(c) shows the situation for a LSW with a uniform interlayer electric field; surprisingly there are three instead of two gapless modes per valley in this case.
We investigate this problem further by studying the dependence on the tunneling amplitude across the LSW shown as the dashed lines in Fig. 5. Without tunneling (Fig. 7(a)), the boundary mode spectrum yields two copies of the gapped BLG spectrum shown in Fig. 6(a), and thus there are two chiral gapless modes in each valley as anticipated by the continuum model. The flat bands represent the states localized on the grain boundary lines and in Fig. 5. The leading effect of turning on the tunneling is that the pair of degenerate flat bands (magenta bands in Fig. 7) are split and become dispersive, as described in Fig. 7(b) and (c). When the tunneling is larger than the electric field induced gap, the flat band split downward is pushed down to the valence band and becomes the third gapless mode shown in Fig. 6(c).
The other two gapless modes (green bands in Fig. 7) localized on the grain boundary of the broken layer are almost degenerate due to the inversion symmetry between the lines and in Fig. 5. This degeneracy is exact at and can be lifted by breaking the inversion symmetry between the left and right domains. We further find that a local potential on or can raise and lower the energies of the green bands. Similarly, a line potential on or can change the energies of the magenta bands. In view of these results we propose a criterion controlled by a hierarchy of energy scales to determine the number of fragile gapless modes in the atomically abrupt LSW shown in Fig. 5:
where is half size of the field-induced gap which saturates if exceeds a critical value ZhangABC (). Two gapless channels emerge if Eq. (12) is satisfied, but extra gapless channels can also appear from the flat bands if Eq. (13) is fulfilled.
When the LSW is made smooth in the sense that it does not produce sufficiently strong intervalley coupling, the tunneling amplitudes near the domain wall in both layers are almost the same as the pristine ones. In such a case, the tunneling between the left and right domains in the broken (continuous) layer would strongly split the two green (magenta) bands at . As a result, only one green and one magenta bands in Fig. 7 survive in the band gap, recovering our earlier continuum results.
We conclude that the gapless interface modes at a LSW are topologically stable only if the potential difference between layers is the dominant energy scale, so that valley is approximately a good quantum number. In the general case the number of domain wall modes can be any integer from to depending on the criteria like that implied by Eq. (12) and (13). The valley-projected topological-state physics of BLG is illustrative of similar physics which occurs in all multi-layer graphene systems ZhangABC (); SQH () and is sensitive to stacking order, and to perpendicular electric fields.
Note added.— After the finalization of this work, a complementary preprint Kim (), which covers closely related material, has appeared.
This work is supported by DARPA under grant SPAWAR N66001-11-1-4110, by the Department of Energy, Office of Basic Energy Sciences under contract DE-FG02-ER45118, and by the Welch Foundation grant TBF1473.
- (1) H. Min, A. H. MacDonald, Phys. Rev. B 77, 155416 (2008).
- (2) E. McCann and V. I. Fal’ko, Phys. Rev. Lett 96, 086805 (2006).
- (3) T. Ohta, A. Bostwick, T. Seyller, K. Horn, and E. Rotenberg, Science 313, 951 (2006).
- (4) E. V. Castro, K. S. Novoselov, S. V. Morozov, N. M. R. Peres, J. M. B. Lopes dos Santos, J. Nilsson, F. Guinea, A. K. Geim and A. H. Castro Neto, Phys. Rev. Lett. 99, 216802 (2007).
- (5) F. Zhang, B. Sahu, H. Min and A. H. MacDonald, Phys. Rev. B 82, 035409 (2010).
- (6) B. E. Feldman, J. Martin and A. Yacoby, Nature Physics 5, 889 (2009).
- (7) J. Martin, B. E. Feldman, R. T. Weitz, M. T. Allen and A. Yacoby, Phys. Rev. Lett. 105, 256806 (2010).
- (8) E. A. Henriksen and J. P. Eisenstein, Phys. Rev. B 82, 041412(R), (2010).
- (9) J. Velasco, L.Jing, W. Bao, Y. Lee, P. Kratz, V. Aji, M. Bockrath, C. N. Lau, C. Varma, R. Stillwell, D. Smirnov, F. Zhang, J. Jung, and A. H. MacDonald, Nature Nano. 7, 156 (2012).
- (10) W. Bao, J. Velasco, F. Zhang, L. Jing, B. Standley, D. Smirnov, M. Bockrath, A. H. MacDonald, and C. N. Lau, Proc. Nat. Acad. Sci. 109, 10802 (2012).
- (11) I. Martin, Y. M. Blanter, and A. F. Morpurgo, Phys. Rev. Lett. 100, 036804 (2008).
- (12) F. Zhang, J. Jung, G. A. Fiete, Q. Niu, and A.H. MacDonald, Phys. Rev. Lett. 106, 156801 (2011).
- (13) D. Xiao, M. C. Chang, and Q. Niu, Rev. Mod. Phys. 82, 1959 (2010).
- (14) J. Li, A. F. Morpurgo, M. Büttiker, and I. Martin, Phys. Rev. B 82, 245404 (2010).
- (15) J. Li, I. Martin, M. Buttiker, and A. F. Morpurgo, Nat. Phys. 7, 38 (2011).
- (16) Z. Qiao, J. Jung, Q. Niu, and A. H. MacDonald, Nano Lett. 11, 3453 (2011).
- (17) J. Jung, F. Zhang, Z. Qiao, and A. H. MacDonald, Phys. Rev. B 84, 075418 (2011).
- (18) A. Vaezi, Y. Liang, D. H. Ngai, L. Yang, and Eun-Ah Kim, arXiv:1301.1690 (2013).