Abstract
Spindependent energy bands and transport properties of ferromagneticstrain graphene superlattices are studied. The high spin polarization appears at the Dirac points due to the presence of spindependent Dirac points in the energy band structure. A gap can be induced in the vicinity of Dirac points by strain and the width of the gap is enlarged with increasing strain strength, which is beneficial for enhancing spin polarization. Moreover, a full spin polarization can be achieved at large strain strength. The position and number of the Dirac points corresponding to high spin polarization can be effectively manipulated with barrier width, well width and effective exchange field, which reveals a remarkable tunability on the wavevector filtering behavior.
Introduction
Graphene has attracted enormous attention from experimentalists and theorists since its discovery. In particular, the high carrier mobility and small spinorbit coupling in graphene make it very promising for applications in nanoelectronics and spintronics. Recently, it is theoretically predicted that^{1,2} depositing a ferromagnetic insulator (FI) such as EuO on graphene can induce an exchange proximity interaction^{1,3}, and the exchange proximity interaction can be treated as an effective exchange field (EEF). The deposition of EuO on graphene has been experimentally realized and its proximity induced ferromagnetization has been confirmed^{4}. Many theoretical works on spin transport through ferromagnetic graphene suggest that the spin current can be controlled by gate voltages^{1,2,5}, magnetic barriers^{6,7}, and local strain^{8,9,10}. Particularly, Dell’Anna found that an inhomogeneous perpendicular magnetic field together with a strong inplane spin splitting can produce a wavevectordependent spinfiltering effect^{6}. Zhai showed that ferromagnetic graphene junctions with a modulated substrate strain can achieve a straintunable spin current^{8}. Recently, Wu has examined that the ferromagnetic graphene system combined with strain or Rashba spinorbit coupling, or both, can induce a spin band gap and achieve complete spin polarization^{10}.
At the same time, graphene superlattices with electrostatic potential or magnetic barrier have also received broad theoretical and experimental investigations^{11,12,13,14,15,16}. In electrostatic potential graphene superlattices, a new Dirac point appears in the band structures^{13,14} and it’s exactly located at zeroaveraged wave number (zero\(\bar{k}\))^{15}. The zero\(\bar{k}\) gap associated with this new Dirac point is insensitive to both lattice constant and structural disorder, resulting in more controllable electronic transport in graphene superlattices. Extra Dirac points in the band structure at zero\(\bar{k}\) have been experimentally observed^{14,17,18}. As a comparison, in magnetic graphene superlattices, new finiteenergy Dirac points are generated in the band structure and the Fermi velocity at zero energy Dirac points is isotropically renormalized^{19,20,21}. Recently, resonant tunneling in ferromagnetic graphene superlattices has been studied and its splitting in the transmission gap can be used to generate an efficient wavevector filter^{22}. However, ferromagnetic graphene superlattices alone cannot suppress the spindependant Klein tunneling^{23}, which results in finite spin polarization.
In addition, the pseudo magnetic field induced by the strain is an efficient method to suppress Klein tunneling^{10,23}, and a local strain can be achieved by patterning grooves, creases, steps, or wells in the substrate where graphene rests^{24,25,26}, so that different regions of the substrate interact differently with the graphene sheet, generating different strain profiles^{27}. Evidence for straininduced spatial modulations in the local conductance of graphene on SiO_{2} substrates has already been reported in experiment^{28}. Building from these literature works, we now consider a ferromagneticstrain graphene superlattice, where the spindependant Klein tunneling is violated. We first discuss the zero\(\bar{k}\) and finiteenergy Dirac points’ locations in the spectrum of a ferromagnetic graphene superlattices in detail. When strain is also considered, we observe that a band gap is induced in the vicinity of finiteenergy Dirac points, and the band gaps for spinup and spindown electrons are present in different energy regions. The spindependent band structure is clearly reflected in the transport properties, which provides a guide for enhancing the spin polarization. The position and number of Dirac points, and the corresponding high spin polarization, can be effectively manipulated by adjusting the barrier width, well width and EEF strength, which demonstrates remarkable tunability on the wavevector filtering behavior.
The paper is organized as follows. In Sec. II, we present the theoretical formalisms and the dispersion relations. The numerical results on band structures and transmission for different spins are shown in Secs. III. Finally, we draw conclusions in Sec. IV.
Computational Models and Methods
Let us consider a onedimensional ferromagneticstrained superlattice in graphene formed by a series of EEF barriers and strained barriers. In our case, we consider a series of FI strips with zaxis magnetizations deposited periodically on the top of graphene to induce the EEF barriers^{1,3}. It has been demonstrated that the EEF between electrons in graphene and localized electrons in an adjacent FI layer is about 5 meV^{1} and can be further enhanced by applying an external electric field perpendicular to the graphene sheet^{3}. In this paper, the local strains are assumed inside these FI stripe regions, which can be induced by a tension along the y direction applied on the substrate rather than the graphene. It is known that graphene can sustain elastic up to 25%^{29,30}. The elastic deformation can be treated as a perturbation to the hopping amplitudes and acts as a pseudogauge potential A _{ s }(r)^{31,32}. Here the pseudogauge potential is induced by the uniaxial strain. As a corollary, the pseudogauge potential is a finite and constant, which is defined as As(r) = As(x) = tβε(1 + σ)^{33}, and σ = 0.165 is the Poisson’s ratio of graphite, t is the nearestneighbour hopping parameter, and ε is the tensile strain. The constant β = ∂lnt/∂lnδ, where δ is the distance between nearest carbon atoms. Several units of such structures are depicted in Fig. 1, and the length of each unit is L = d _{1} + d _{2}. The lowenergy effective Hamiltonian for ferromagneticstrain graphene can be written as
where v _{ F } ≈ 10^{6} m/s is the Fermi velocity, τ _{ z } = ±1 for K and K ^{'} valleys, σ _{ i } and s _{ i } (i = x, y, z) are the Pauli matrices acting on the sublattice (A, B) and physical spin (↑, ↓) spaces, respectively. Due to the translational invariance in the y direction, the wave function in the j th ferromagneticstrained barrier can be presented as \(\tilde{{\rm{\Psi }}}={\rm{\Psi }}(x){e}^{i{k}_{y}x}\) with
where \({k}_{j}=\tfrac{(Es{M}_{j})}{(\hbar {v}_{F})}\) (s = ±1 for spinup and spindown electrons) and k _{ yj } = k _{ y } + τ _{ z }(A _{ sj })/(ħv _{ F }). k _{ y } = (Esinθ _{0})/(ħv _{ F }), θ _{0} is the incident angle. q _{ j } = sign (k _{ j }) (k _{ j } ^{2} − k _{ yj } ^{2})^{1/2} for k _{ j } ^{2} > k _{ yj } ^{2}, otherwise q _{ j } = sign(k _{ j }) i(k _{ yj } ^{2} − k _{ j } ^{2})^{1/2}, and a _{ j }(b _{ j }) is the amplitude of the forward (backward) propagating wave. For the well region, the above equations are still valid and only require that M _{ j } = 0 and A _{ sj } = 0. Inside the same barrier or well region, the wave functions at any two positions x and x + Δx can be related via the transfer matrix^{16}
with θ _{j} = arcsin(k_{yj}/k_{j}). Furthermore, the overall Tmatrix for the N regions is simply a product of matrices:
Here the w _{ j } is the width of the jth potential region. And we can connect the input and output wave functions by the relation: Ψ(x _{ N }) = XΨ(x _{0}), where the Ψ(x _{ N }) and Ψ(x _{0}) can be written as:^{15}
and
Here, θ _{ N } is the exit angle at the exit end, Ψ_{ i }(E, k _{ y }) is the incident wave packet of the electron, \({r}_{s,{\tau }_{z}}\) is the spin/valley resolved reflection coefficient and t _{s,τz} is the spin/valley resolved transmission coefficient, respectively. Solving the above two equations, we find the r _{s,τz} and t _{s,τz} can be given
Once the transmission coefficient is obtained, the spin/valley resolved conductance G _{ s,τz } of the system at zero temperature is written as G _{s,τz} = G _{0}∫_{0} ^{(π)/(2)} T _{s,τz} cosθ _{0} dθ _{0}, where T _{s,τz} = t _{s,τz} ^{2}, G _{0} = 2e ^{2} mv _{ F } L _{ y }/ħ ^{2} and L _{ y } is the width of the graphene stripe in the y direction. Meanwhile, the spin polarizations are defined as
Results and Discussion
In order to understand the transport properties, it is instructive to first investigate the electronic band structure for the ferromagneticstrain graphene supperlatice. According to the Bloch’s theorem^{15}, the electronic dispersion for any transversal wave number follows the relation:
Here K _{ s,τz } is Bloch wave vector. Γ_{1} and Γ_{2} are the transfer matrixes for one barrier and one well, respectively. q _{1} = sign(E − sM)\(\sqrt{{(\tfrac{E{}_{s}M}{\hslash {v}_{F}})}^{2}\,\,{({k}_{y}+{\tau }_{z}\tfrac{{A}_{s}}{\hslash {v}_{F}})}^{2}}\) for \({(\frac{EsM}{\hslash vF})}^{2} > {({k}^{y}+{\tau }_{z}\tfrac{As}{\hslash {v}_{F}})}^{2}\), otherwise \({q}_{1}={\rm{sign}}(EsM)i\) \(\sqrt{{({k}_{y}+{\tau }_{z}\tfrac{{{\rm{A}}}_{s}}{\hslash {v}_{F}})}^{2}{(\tfrac{EsM}{\hslash {v}_{F}})}^{2}}\). \({q}_{2}={\rm{sign}}(E)\sqrt{\tfrac{{E}^{2}}{{(\hslash {v}_{F})}^{2}}+{k}_{y}^{2}}\,{\rm{for}}\,\tfrac{{E}^{2}}{{(\hslash {v}_{F})}^{2}} > {{k}_{y}}^{2}\), otherwise \({q}_{2}={\rm{sign}}(E)i\sqrt{{k}_{y}^{2}\tfrac{{E}^{2}}{{(\hslash {v}_{F})}^{2}}}\). Using cos(K _{s,τz} L) ≤ 1, we can find the real solution of \({K}_{s,{\tau }_{z}}\) for passing band_{ s }. Otherwise, the nonexistence of real \({K}_{s,{\tau }_{z}}\) indicates a band gap^{34}.
Now let us use the above equations to calculate the electronic band structures under different strain strength. The transmissions of electrons in K and K ^{'} valleys show mirror symmetry^{10,23}, so we focus only on the spin transport for the valley K. When A _{ s } = 0 (Fig. 2a,d), we find that the zero\(\bar{k}\) Dirac point is given at \(E=\frac{sM}{2}\), k _{ y } = 0. When strain is considered (Fig. 2b,c,e,f), we find that the zero\(\bar{k}\) Dirac point is shifted to \({k}_{y}=\frac{{A}_{s}}{2\hslash {v}_{F}}\) along \(E=\frac{sM}{2}\). Here \(\hslash {v}_{F}=(\frac{4.14\times {10}^{12}}{2\pi }\,{\rm{meV}}\cdot {\rm{s}})\,({10}^{15}\,{\rm{nm}}/{\rm{s}})=\frac{2070}{\pi }\,{\rm{meV}}\cdot {\rm{nm}}\). In other words, the Dirac point shifts to \({k}_{y}=\frac{{A}_{s}}{2\hslash {v}_{F}}\) in k space, but the energy is invariable. Such a result can be solved by the dispersion relation of Eq. (6).
Applying the implicit function theorem, the gradient of the dispersion relation will be zero only if sin(q _{1} d _{1}) = sin(q _{2} d _{2}) = 0 and cos(q _{1} d _{1}) = cos(q _{2} d _{2}) = 1. When A _{ s } = 0, the following equations are satisfied
Here m, n are integers. The equation (6) shows that cos(K _{s,τz} L) = cos(q _{1} d _{1} ± q _{2} d _{2}) for k _{ y } = 0 and A _{ s } = 0, which indicates that K _{s,τz} always has real solutions for any E and M; that is, the location of the crossing point of the bands exactly appears at k _{ y } = 0. So under the condition of k _{ y } = 0 and A _{ s } = 0, one can get \(E=\tfrac{sM{d}_{1}}{{d}_{1}+{d}_{2}}+\hslash {v}_{F}{\rm{\pi }}\tfrac{m+n}{{d}_{1}+{d}_{2}}\). If the condition q _{1} d _{1} =− q _{2} d _{2} = mπ is satisfied, the solution is \(E=\tfrac{sM{d}_{1}}{{d}_{1}+{d}_{2}}\), which is the socalled zero\(\bar{k}\) Dirac point^{15}. Further when d _{1} = d _{2}, the zero\(\bar{k}\) Dirac point is located at \(E=\tfrac{sM}{2}\), k _{ y } = 0. This way, we can also locate the other crossing points corresponding to the finiteenergy Dirac points^{20}, at \(E=\tfrac{sM{d}_{1}}{{d}_{1}+{d}_{2}}\) + \(\hslash {v}_{F}\pi \frac{l}{{d}_{1}+{d}_{2}}\)(l = ±1, ±2 \(\cdots \)) and k _{ y } = 0. From Fig. 2a,d, one can observe that the finiteenergy Dirac points for the spinup (spindown) band is exactly located in E = 71.15 meV, −31.15 meV, −81.15 meV (81.15 meV, 31.15 meV, −71.15 meV). Moreover, Fig. 2a,d also suggest that the spinup and spindown Dirac points don’t always coincide, which plays a key role in spindependent transport, but it is noted that the bands always cross at k _{ y } = 0.
In addition, if A _{ s } ≠ 0, one can find
When q _{1} d _{1} =− q _{2} d _{2} = mπ is also satisfied, we can get
The equation (9) is tenable under the conditions Ed _{1} − sMd _{1} =−Ed _{2} and k _{ y } d _{1} + \(\frac{{A}_{s}}{\hslash {v}_{F}}\) d _{1} = −k _{ y } d _{2}, so E = \(\frac{sM{d}_{1}}{{d}_{1}+{d}_{2}}\), \({k}_{y}=\tfrac{As{d}_{1}}{\hslash {v}_{F}({d}_{1}+{d}_{2})}\) is one solution of Eq. (9), which corresponds to the zero\(\bar{k}\) Dirac point. If d _{1} = d _{2}, zero\(\bar{k}\) Dirac point is located at \(E=\tfrac{sM}{2}\), \({k}_{y}=\tfrac{As}{2\hslash {v}_{F}}\). However, it is more difficult to find analytic solutions of the finiteenergy Dirac points like the analytic results obtained by solving Eq. (7). But numerical calculations show that the finiteenergy Dirac points are strongly affected by the strain strength. Due to the effects of strain, the finiteenergy Dirac points are not only shifted in energy but also decreased in number (Fig. 2b,e), even disappear completely for large strain strengths (Fig. 2c,f). Then, there emerges an energy gap in the vicinity of the vanished finiteenergy Dirac points with further increasing the strain strength. And the energy gaps for the spinup and spindown bands don’t fully overlap. These characters mean that the increasing of A _{ s } may be used to enhance the spin polarization in ferromagneticstrain graphene superlattices.
The above discussions on the band structures should be helpful for understanding the spindependent transport. Figure 3 displays the spindependent transmission T _{ s }, spindependent conductance G _{ s } and spin polarization along z direction P _{ z } of the ferromagneticstrain graphene superlattices under different strain. Here we only consider A _{ s } = 0 and A _{ s } = 60 meV, and take the superlattice period number n = 10. In the absence of strain (Fig. 3a,b), the transmission shows a spindependent Klein tunneling and embodies the mirror symmetry about θ = 0. But the transmission for upspins is different from that for downspins, especially at the locations of the Dirac points where the transport channels for up(down) spins are finite, while the transport channels for down(up) spins are large. These characters ensure that the two spin conductance channels are obviously different at these Dirac points (as seen in Fig. 3e), and finite spin polarization appears (as seen in Fig. 3(g)). When strain is considered (Fig. 3c,d), we find that the mirror symmetry with θ = 0 is destroyed because of the shifted Dirac points by the strain, and the spindependent Klein tunneling is suppressed due to the spindependent band gap induced by the strain. It is noted that the spindependent transmission gaps also induce zero\(\bar{k}\) Dirac points nearby because the spindependent waves inside the potential barrier are evanescent waves when the relation \({(EsM)}^{2} < {({k}_{y}+{\tau }_{z}{A}_{s})}^{2}\) is satisfied. Then we find that the spinup and spindown conductances are totally different around those disappeared Dirac points. Especially, in the vicinity of E = 20 meV, 71.15 meV (E =−20 meV, −71.15 meV), the spindepended conductance G _{↓} (G _{↑}) shows a broad peak, while G _{↑} (G _{↓}) approaches zero (Fig. 3f), so fully spin polarized plateaus with large spinpolarized currents are achieved around these Fermi energies (as seen in Fig. 3g). In addition, spin polarization oscillations are obtained as seen in Fig. 3g, which can be used as a spin switch by modulating the Fermi energy.
The above discussions show that high spin polarizations always appear in the vicinity of the vanished Dirac points. And the positions of Dirac points in (E, k _{ y }) space can be controlled by the barrier and well widths. Figure 4a–c show the band structures with different barrier and well widths. The locations of the zero\(\bar{k}\) Dirac points move towards E = 0 and k _{ y } = 0 with gradually reduced d _{1}/(d _{1} + d _{2}) ratio at fixed heights of potentials. The locations of the band gaps around the vanished finiteenergy Dirac points move toward E = 0 too. In addition, the number of band gaps increases with the increase of the lattice constant d _{1} + d _{2}. The reason of that is the zero\(\bar{k}\) Dirac points is exactly located \(E=\tfrac{sM{d}_{1}}{{d}_{1}+{d}_{2}}\), \({k}_{y}=\tfrac{As{d}_{1}}{{d}_{1}+{d}_{2}}\), which is determined by the d _{1}/(d _{1} + d _{2}) ratio. The finiteenergy Dirac points are located at \(E=\tfrac{sM{d}_{1}}{{d}_{1}+{d}_{2}}+\hslash {v}_{F}{\rm{\pi }}\tfrac{l}{{d}_{1}+{d}_{2}}\), which depends not only on the d _{1}/(d _{1} + d _{2}) ratio but also the lattice constant d _{1} + d _{2}. So we can modulate the location and number of high spin polarization regions by adjusting the the d _{1}/(d _{1} + d _{2}) ratio and the lattice constant.
Next we consider the spindependent conductance G _{ s } (Fig. 4d–f) and spin polarization P _{ z } (Fig. 4h–r) of an electron passing through the ferromagneticstrain graphene superlattices with different width. Comparison between Fig. 4a–f indicates that the distribution of transmission spectra is completely consistent with the band structures, that is, strong transmission regions correspond to the transmission bands and forbidden transmission regions correspond to the band gaps. Then, the location of high spin polarization approaches E = 0 with the decrease of d _{1}/(d _{1} + d _{2}) ratio, and the number of high spin polarization regions increases with increasing lattice constants (Fig. 4h–r). Therefore the increase of lattice constants makes the spin polarization oscillations more obvious.
In addition, the height of potentials can also affect the locations of the Dirac points. Figure 5a shows the spin polarization with respect to M and E for d _{1} = d _{2} = 20 nm. In the absence of EEF, the spin polarization is zero (Fig. 5a) due to the spin degeneracy (as seen in Fig. 5b). And the spin polarization initially increases and then decreases with increasing the EEF strength for M ≥ A _{ s } (Fig. 5a). The reason is that when the EEF strength M ≥ A _{ s }, the gaps around the Dirac points are finite (as seen in Fig. 5c,d) and the crossing points even reappear for larger M (as seen in Fig. 5e,f), which leads to both upspins and downspins having transport channels around the Dirac points therefore the spin polarization is reduced. So too large M does not guarantee effective spin filtering in such ferromagneticstrain graphene superlattices. We also find that the high spin polarization regions are shifted away from zero energy owing to the shift of Dirac points away from E = 0 with the increasing of EEF.
Summary
In summary, we studied the spindependent band structures and transport properties of graphene under a periodic effective exchange field and strain, where the spindependant Klein tunneling is disrupted. We discussed the zero wave number Dirac points’ and finiteenergy Dirac points’ locations on the spectra of ferromagnetic graphene superlattices in detail. The spinup and spindown Dirac points are present on the energy spectra alternately, which results in finite spin polarization. When strain is considered, band gaps are induced around the finiteenergy Dirac points, and high spin polarization is achieved in the vicinity of these Dirac points. The position, and number of the Dirac points can be effectively manipulated by adjusting the barrier width, well width and EEF strength, which leads to tunable spin polarization. We hope these results are helpful for understanding the electronic properties for spin transports and can offer guidance to potential applications of the spin filtering devices.
References
 1.
Haugen, H., HuertasHernando, D. & Brataas, A. Spin transport in proximityinduced ferromagnetic graphene. Phys. Rev. B 77, 115406 (2008).
 2.
Yang, H. X. et al. Proximity effects induced in graphene by magnetic insulators: Firstprinciples calculations on spin filtering and exchangesplitting gaps. Phys. Rev. Lett. 110, 046603 (2013).
 3.
Semenov, Y., Kim, K. & Zavada, J. Spin field effect transistor with a graphene channel. Appl. Phys. Lett. 91, 153105 (2007).
 4.
Swartz, A. G., Odenthal, P. M., Hao, Y., Ruoff, R. S. & Kawakami, R. K. Integration of the ferromagnetic insulator euo onto graphene. ACS Nano 6, 10063–10069 (2012).
 5.
Yokoyama, T. Controllable spin transport in ferromagnetic graphene junctions. Phys. Rev. B 77, 073413 (2008).
 6.
Dell’Anna, L. & De Martino, A. Wavevectordependent spin filtering and spin transport through magnetic barriers in graphene. Phys. Rev. B 80, 155416 (2009).
 7.
Khodas, M., Zaliznyak, I. A. & Kharzeev, D. E. Spinpolarized transport through a domain wall in magnetized graphene. Phys. Rev. B 80, 125428 (2009).
 8.
Zhai, F. & Yang, L. Straintunable spin transport in ferromagnetic graphene junctions. Appl. Phys. Lett. 98, 062101 (2011).
 9.
Niu, Z. Spin and valley dependent electronic transport in strain engineered graphene. J. Appl. Phys. 111, 103712 (2012).
 10.
Wu, Q.P., Liu, Z.F., Chen, A.X., Xiao, X.B. & Liu, Z.M. Generation of full polarization in ferromagnetic graphene with spin energy gap. Appl. Phys. Lett. 105, 252402 (2014).
 11.
Marchini, S., Günther, S. & Wintterlin, J. Scanning tunneling microscopy of graphene on ru(0001). Phys. Rev. B 76, 075429 (2007).
 12.
Vázquez de Parga, A. L. et al. Periodically rippled graphene: Growth and spatially resolved electronic structure. Phys. Rev. Lett. 100, 056807 (2008).
 13.
Barbier, M., Vasilopoulos, P. & Peeters, F. M. Dirac electrons in a kronigpenney potential: Dispersion relation and transmission periodic in the strength of the barriers. Phys. Rev. B 80, 205415 (2009).
 14.
Brey, L. & Fertig, H. A. Emerging zero modes for graphene in a periodic potential. Phys. Rev. Lett. 103, 046809 (2009).
 15.
Wang, L.G. & Zhu, S.Y. Electronic band gaps and transport properties in graphene superlattices with onedimensional periodic potentials of square barriers. Phys. Rev. B 81, 205444 (2010).
 16.
Zhao, P.L. & Chen, X. Electronic band gap and transport in fibonacci quasiperiodic graphene superlattice. Appl. Phys. Lett. 99, 182108 (2011).
 17.
Park, C.H., Yang, L., Son, Y.W., Cohen, M. L. & Louie, S. G. Anisotropic behaviours of massless dirac fermions in graphene under periodic potentials. Nat. Phys. 4, 213–217 (2008).
 18.
Yankowitz, M. et al. Emergence of superlattice dirac points in graphene on hexagonal boron nitride. Nat. Phys. 8, 382–386 (2012).
 19.
Tan, L. Z., Park, C.H. & Louie, S. G. Graphene dirac fermions in onedimensional inhomogeneous field profiles: Transforming magnetic to electric field. Phys. Rev. B 81, 195426 (2010).
 20.
Dell’Anna, L. & De Martino, A. Magnetic superlattice and finiteenergy dirac points in graphene. Phys. Rev. B 83, 155449 (2011).
 21.
Le, V. Q., Pham, C. H. & Nguyen, V. L. Magnetic kronig–penneytype graphene superlattices: finite energy dirac points with anisotropic velocity renormalization. J. Phys.: Condens. Matter 24, 345502 (2012).
 22.
Lu, W.T., Li, W., Wang, Y.L., Jiang, H. & Xu, C.T. Tunable wavevector and spin filtering in graphene induced by resonant tunneling. Appl. Phys. Lett. 103, 062108 (2013).
 23.
Zhang, Y.T. & Zhai, F. Strain enhanced spin polarization in graphene with rashba spinorbit coupling and exchange effects. J. Appl. Phys. 111, 033705 (2012).
 24.
Li, X., Wang, X., Zhang, L., Lee, S. & Dai, H. Chemically derived, ultrasmooth graphene nanoribbon semiconductors. Science 319, 1229–1232 (2008).
 25.
Low, T. & Guinea, F. Straininduced pseudomagnetic field for novel graphene electronics. Nano Lett. 10, 3551–3554 (2010).
 26.
Lu, J., Neto, A. C. & Loh, K. P. Transforming moiré blisters into geometric graphene nanobubbles. Nat. Commun. 3, 823 (2012).
 27.
Pereira, V. M. & Castro Neto, A. H. Strain engineering of graphene’s electronic structure. Phys. Rev. Lett. 103, 046801 (2009).
 28.
Teague, M. et al. Evidence for straininduced local conductance modulations in singlelayer graphene on sio2. Nano Lett. 9, 2542–2546 (2009).
 29.
Lee, C., Wei, X., Kysar, J. W. & Hone, J. Measurement of the elastic properties and intrinsic strength of monolayer graphene. Science 321, 385–388 (2008).
 30.
Kim, K. S. et al. Largescale pattern growth of graphene films for stretchable transparent electrodes. Nature 457, 706–710 (2009).
 31.
Pereira, V. M., Castro Neto, A. H. & Peres, N. M. R. Tightbinding approach to uniaxial strain in graphene. Phys. Rev. B 80, 045401 (2009).
 32.
Fujita, T., Jalil, M. & Tan, S. Valley filter in strain engineered graphene. Appl. Phys. Lett. 97, 043508 (2010).
 33.
Liu, Z.F. et al. Helical edge states and edgestate transport in strained armchair graphene nanoribbons. Sci. Rep. 7, 8854 (2017).
 34.
Fan, X., Huang, W., Ma, T., Wang, L.G. & Lin, H.Q. Electronic band gaps and transport properties in periodically alternating monoand bilayer graphene superlattices. EPL 112, 58003 (2015).
Acknowledgements
This work is supported by the National Natural Science Foundation of China (Grant Nos. 11764013, 11364019, 11464011, 11365009 and 11664019), the China Scholarship Council (File NOs 201509795008 and 201509795010), the Science Foundation for Distinguished Young Scholars in Jiangxi Province (Grant No. 2016BCB23032) and NSERC Discovery grant RGPIN418415 and RGPIN04178, and the Canada First Research Excellence Fund.
Author information
Affiliations
Contributions
Q.P.W. conceived the idea and performed the calculation. Z.F.L., Q.P.W. and G.X.M. contributed to the interpretation of the results and wrote the manuscript. A.X.C. and X.B.X. contributed in the discussion. All authors reviewed the manuscript.
Corresponding authors
Correspondence to ZhengFang Liu or GuoXing Miao.
Ethics declarations
Competing Interests
The authors declare that they have no competing interests.
Additional information
Publisher's note: Springer Nature remains neutral with regard to jurisdictional claims in published maps and institutional affiliations.
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
Received
Accepted
Published
DOI
Further reading

Valleydependent transport properties of electrons in a graphene with magnetic field and strained barrier
Journal of Magnetism and Magnetic Materials (2019)

Effect of strained barrier on valley polarization in magneticmodulated graphene
International Journal of Modern Physics B (2019)
Comments
By submitting a comment you agree to abide by our Terms and Community Guidelines. If you find something abusive or that does not comply with our terms or guidelines please flag it as inappropriate.