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The rather large amount of nitrogen correspond to results from Rutherford backscattering experiments which yielded a ratio of about 2:1 of sulfur to nitrogen 12 and are also in agreement with previously reported values. 6 A number of new bands appear in the calculated spectra from the nitrogen containing clusters. However, while the clusters without nitrogen lead to very similar spectra, the modified systems show pronounced differences. In general the spectra 2-4 do not compare well with the experimental data, while the calculated spectrum 1 is surprisingly close to the experimental data. In particular, the broad shoulder of the main band is here well reproduced. The additional band at 1080 cm -1 is still present, but does not relate to nitrogen incorporation as discussed above. | cond-mat.mtrl-sci |
With this in mind, a specific feature that we focus upon in our analysis is the presence of temperature induced neutron scattering intensity at low frequencies much smaller than the zero temperature gap. The origin of this intensity is intraband scattering. The analogous phenomenon in Ising-like antiferromagnetic spin chains was first pointed out by J. Villain 28 and was first observed in the anisotropic spin chain material CsCoBr 3 . 29,30 For Ising antiferromagnets, the relevant excitations are domain walls in the anti-ferromagnetic order. In contrast, in the dimer model the relevant excitations correspond to low lying magnons. | cond-mat.str-el |
The spinon excitations of individual chains are confined by even an infinitesimal coupling strength J ⊥ [18,21]. The two S = 1 2 spinons form a bound state giving rise to singlet and triplet excitation branches. In the case of the strongly coupled ladder (J ⊥ >> J || ), the elementary excitation is a triplet. Lake et al. [21] carried out neutron scattering experiments on the weakly-coupled (J ⊥ << J || ) ladder material CaCu2O3 and obtained evidence of the singlet excitation mode. The spinon continuum was observed at high energies for which the chains are effectively decoupled. | cond-mat.str-el |
Similar enhancement of the relaxation time τ in with the increase of the magnetic field is found for sample N1 (not shown). CONCLUSION We have studied the nonlinear response of 2D electrons placed in crossed electric and quantized magnetic fields at low temperatures. The resistance of 2D electrons decreases strongly with an increase of the electric field. The decrease of the resistance is in good quantitative agree-ment with theory considering the nonlinear response as a result of non-uniform spectral diffusion of 2D electrons limited by inelastic electron scattering. Comparison between the experiments and the theory has revealed different regimes of the electron inelastic relaxation. At low magnetic fields, at which the Landau levels are well overlapped and the spectral diffusion is weakly modulated with the electron energy, the inelastic scattering rate is found to be proportional to the square of the temperature T 2 in temperature interval (2-10 (K)). | cond-mat.mes-hall |
2 is the calculated critical current I c /S in the limit of λ t → 0. According to Fig. 2 we conclude that the effect of the finite penetration depth λ t is less important to describe the results of Ref. [12]. The reason why both I 0 c and I P c remain finite in the limit of d 1 → 0 is understood as follows. Slonczewski assumed that the transverse spin current injected into the free layer is absorbed at the interface, and thus, STT is independent of the thickness of the free layer. | cond-mat.mes-hall |
[17] and Wang et. al. [18], respectively. The buckling behavior of the nanowires was observed and Young modules were estimated. Branched nanowires of various types offer another approach to increase structural complexity and physical properties [19]. In the paper by Wang et al. | cond-mat.mtrl-sci |
[21] could not observe a clear difference of FSs between CeRu 2 Si 2 and LaRu 2 Si 2 , Yano et al. [22] claimed the observation of 4f -delocalized FSs for CeRu 2 Si 2 and 4f -localized FSs for CeRu 2 Ge 2 even in the paramagnetic state. Therefore, it is expected that soft Xray ARPES experiments for CeRu 2 (Si 1-x Ge x ) 2 can verify the existence of the FS crossover around the QCP of this system. In order to enhance the photoemission signals of the Ce 4f electrons, we have performed ARPES experiment in the Ce 3d→4f resonance energy region and successfully observed the strongly c-f hybridized FS both for CeRu 2 Si 2 and CeRu 2 (Si 0.82 Ge 0.18 ) 2 , below and above x c = 0.07, respectively. ARPES was also measured for LaRu 2 Si 2 as a reference material where the 4f electrons do not participate in the FS formation. All the measured samples were single crystals grown in the procedures described in Ref. | cond-mat.str-el |
The latter condition ensures that the partial charges are larger in the center of the construction cell and smaller on its edges, and hence that this cell is close to the original one. We did not find a way to derive the solution of this problem in a general way, for any value of l. We thus determined the solution for fixed values of l up to l = 6. In all these cases the first λ p function presents the same shape :λ 1 (x) = x l-1 (l -1)! (38) We reasonably assume that this expression is valid for any value of l. As we will see, it is possible to show, a posteriori, that the λ p functions fulfill the first three conditions. We will now determine the function λ p for any l, using the above expression of λ 1 and relation 37. In equation 37, the m k,i constants can be replaced by the value of the moments obtained for x = 0 : l p=1 λ p (x) x + p -1 - l 2 k = l p=1 λ p (0) p -1 - l 2 k (39) with 0 ≤ k ≤ l -1. | cond-mat.str-el |
(9) On the basis of the diagrams depicted on the Fig. 2, instead it is possible to establish the following Dyson's type equation for g σσ ′ : g σσ ′ (τ -τ ′ |λ) = Λ σσ ′ (τ -τ ′ |λ) + σ1σ2 β 0 dτ 1 β 0 dτ 2 Λ σσ1 (τ -τ 1 )λG 0 σ1σ2 (τ 1 -τ 2 )λg σ2σ ′ (τ 2 -τ ′ ),(10)where Λ σσ ′ (τ -τ ′ |λ) = g 0 σσ ′ (τ -τ ′ ) + Z σσ ′ (τ -τ ′ |λ). (11) Here Z σσ ′ is the new correlation function which contains an infinite sum of the irreducible Green's functions. As it was underlined above this function contains all spin, charge and pairing fluctuations and is the main element of our diagram technique. Diagram representation of the correlation function Λ σσ ′ (τ -τ ′ |λ) is depicted on the Fig. 3 The series expansion of the full propagator G σσ ′ (τ -τ ′ ) can gives us more detailed representation of this quantity. | cond-mat.str-el |
This value is approximately half of the expected value considering ferrimagnetic order. The saturation magnetization per mole of GdCa 1/3 MO expected from the Gd 3+ s = 7/2, l = -Mn 3+ )+ 1/3x3/2 (from Mn 4+ 7 Considering these magnetizations, we expect a M s 160 emu/cm 3 . 7 Magnetic hysteresis loops also show a high paramagnetic like signal at different temperatures (see figure 3). Although at low temperatures it could be associated with sublattice rotation due to canting, 7 it could also be associated with phase coexistence. It is important to remark that the contribution of paramagnetic Gd moment alone can not explain the high paramagnetic signal, because in this case a high M s should be expected from the non compensated ferromagnetic Mn moments. Figure 5 shows the temperature dependence of the coercive field,H c =| H c1 H c2 | /2 , where H c1 and H c2 are the fields for zero magnetization at both branches of the hysteresis loops for field excursions up to 1 and 3 T. The H c temperature dependence shows a non monotonic decrease when the temperature is raised, resulting quite different to the continuous and smooth decrease measured by O. Peña et al. | cond-mat.mtrl-sci |
Introduction and Motivation Following the publication of Joannopoulos et al (1995) 1 the field of photonic crystals has steadily grown. Parallel to that has been the field of phononic crystal research. Most of the work in both of these areas has been in proof of concept to improve theoretical understanding, or computational demonstration of simple devices. The band-gaps in both types of crystals are controlled by material composition, lattice spacing, crystal-packing arrangement, crystal orientation, and size of the elements comprising the crystal. There have already been far too many papers on phononic crystals to review here, however, of particular interest for acoustic band-gap engineering are those by Garcia et al. (2003), 2 Kushwaha and Halevi (1994), 3 Lai et al. | cond-mat.mtrl-sci |
Each of the two localized bulk quasiparticles carries a zero mode, described by a localized Majorana operator. We denote the two bulk Majorana operators by γ u , γ d , with the subscript indicating the edge to which the quasiparticle couples. The two-dimensional Hilbert space created by the two Majorana modes is spanned by the two eigenvectors of the operator iγ u γ d . To examine the effects of bulk-edge coupling we couple γ u to the upper edge and γ d to the lower edge, both at x = 0. The Lagrangian density for this coupling is L b-e = i λ u ψ u (x)γ u + λ d ψ d (x)γ d δ(x) . (1) The Lagrangian L u m + L d m + L b-e introduces the time scales t λu(d) defined above. | cond-mat.mes-hall |
Of the three, only the first interface is found to be conducting and shows metallic behavior. Furthermore, a minimum thickness of 5 unit cells (uc) of LaVO 3 is required to form a conducting n-type interface. The thickness dependent transport properties and low temperature anomalous Hall effect of the n-type interface suggests coupling of the interface electrons to the Mott insulator LaVO 3 . LaVO 3 thin films were grown on SrTiO 3 substrates by pulsed laser deposition (PLD) using a LaVO 4 polycrystalline target. Most of the structures were grown at 600 o C under an oxygen partial pressure of 1 x 10 -6 Torr, with a laser fluence of 2.5 J/cm 2 , following our previous optimization for two-dimensional layer-by-layer growth of LaVO 3 [15]. By depositing LaVO 3 directly on TiO 2 terminated (001)-oriented SrTiO 3 substrates, the n-type interface 1) is formed, as shown in Fig. | cond-mat.str-el |
16 studied the effect of disorder in graphene by considering the effect of vacancies on the honeycomb lattice. For a finite density of vacancies, they found that the density of states at the Dirac point is zero for the "full Born approximation" (equivalent to our T -matrix approximation) and non-zero for the "full self-consistent Born approximation" (equivalent to our self-consistent Tmatrix approximation). Our results are consistent with theirs, even though the regimes that are studied are different. Vacancies correspond to the limit where the impurity potential U 0 → ∞, whereas this work is more concerned with the weak impurity-scattering limit. Pereira at al. 17 considered, among several different models of disorder, both vacancies and randomness in the on-site energy of the honeycomb lattice. | cond-mat.mes-hall |
The relaxation rate due to staggered fluctuations can be calculated following the prescription of Barzykin 36 . For the purpose of comparison of theory with experiment he defined the normalized dimensionless NMR spinlattice relaxation rate at low-temperature (1/T 1 ) norm = hJ1 A 2 th T1 ≈ 0.3, where A th is A hf (2hγ/2π). Assuming the fluctuations to be correlated because of the exchange J 1 along the chains, 1/T 1 can be written as 1/T 1 = 0.3 hJ1/A 2 th . Using this expression, (1/T 1 ) at the 31 P site was calculated to be about 129 sec, -1 whereas our experimental value is 50 sec, -1 in the 2 K ≤ T ≤ 30 K range. The experimental value is about two times smaller than the theoretical value, likely due to the effect of the geometrical form factor. Further on a weak logarithmic increase in 1/T 1 is theoretically expected at low temperatures. | cond-mat.str-el |
The sample surface shows a 60° section rather than a 90° cross-section. The layers appear therefore thicker, and the axis angles are smaller. Furthermore, since we are looking at a 2D section and not a 3D representation of nacre, the components of all angles towards and away from the reader are not apparent. Both effects contribute to an underestimation of the stacking direction angles. It is possible, therefore, that the real angles at the boundary vary by more than ±60°, even possibly approaching the limit of randomly directed stacks, which vary between ±90°. It is remarkable how, from the complete disorder of random or near-random directions, stacking of co-oriented tablets is harnessed within a mere 50 !m. | cond-mat.mtrl-sci |
. . × ( µ G ) gn(97) This last expression involves explicitly the cumulants defined in Eq. ( 79) and as a result, the order-m cumulant can be expressed as a combination of lower order cumulants: ( ) c m = - m-1 G=1 µ∈S m G [µ] ( µ1 µ2 µ G ) c G(98) The demonstration can be readily extended to the case where the operators in the l.h.s. of Eq. ( 98) are arbitrary con- ) c m = - m-1 G=1 µ∈S Γ G [µ] ( µ1 µ2 µ G ) c G ,(99) which concludes the recursion proof. | cond-mat.str-el |
1A, which eventually translates into a time dynamics of about 500 ns for coherently swapping the two spins. In the geometry shown in Fig. 1B the exchange coupling J can be much larger, on the order of J ≈ 4 • 10 -5 meV for detunigs on the order of ∆ ≈ 4 • 10 -3 meV, which implies a time dynamics of about 20 ns for swapping the two spins coherently. In order to control the exchange coupling J, one should be able in principle to change the Zeeman splitting or the orbital level spacing. In InAs QDs the Zeeman splitting can be changed very efficiently by changing the dot size along the wire direction, 6 in both cases in Fig. 1 Considering the case of two QDs in the cavity, one way to decouple them is by tuning the g-factors so that ∆ 1 = -∆ 2 , as can be seen from Eq. | cond-mat.mes-hall |
This is true, even though in states like the (3,3,1) state where the spin degree of freedom is active and one would expect a 2-channel Kondo model to be realized (because the situation is similar to that discussed in Ref. [96,97]). However, this 2-channel Kondo model will not be "topologically protected" in the way that it is for the Pfaffian. Because of residual Zeeman coupling to the elec-tron's spin in the quantum Hall state the spin up and spin down edge modes will be slightly shifted with respect to one another on the edge. This will lead to different tunneling matrix elements between the edge and dot for different spin orientations, breaking the channel symmetry in the effective 2channel Kondo model. Thus, the low energy fixed point will be described by a single-channel Kondo model, rather than the 2-channel version. | cond-mat.str-el |
The structure of the solution is most easily grasped referring to Fig. 2(a). It is organized in three types of zones: (A), (B), and (D), separated by "frontiers" marked by solid lines. The PS stress reduces to σ PS = g(x)/2 or g(y)/2 in zone (B), to σ PS = 0 in zone (D), and to σ PS = [g(x) + g(y)]/2 in zone (A). It thus vanishes in the square of length 2a consisting in the union of (D) and of the void (V). This is more easily understood in terms of characteristics. | cond-mat.mtrl-sci |
By using positive semidefinite operators ( P ) in deducing exact ground states (|Ψ g ), one first casts the Hamiltonian ( Ĥ) in a positive semidefinite form Ĥ = P + C, where C is a constant depending on the coupling constants of the starting Ĥ. Since zero is the minimum possible eigenvalue of P , the unapproximated ground state is obtained from the requirement P |Ψ g = 0. In the process of transforming the starting Hamiltonian in a positive semidefinite form, mostly one finds or uses P = nmax n Pn , where the number n max of different types of positive semidefinite terms often can reach even n max = 8 -10, their structure being diverse, containing linear, bi-linear, cubic or quartic combinations of the starting fermionic operators 15 . The requirement for the ground state in the presence of different Pn contributions becomes Pn |Ψ g = 0 for all n ∈ [1, n max ], hence the deduced |Ψ g represents separately the ground state of different parts of the Hamiltonian, often being the ground state separately of the kinetic and interacting part of Ĥ, which reduces considerably the application possibilities of the method. In this paper we overcome this inconvenience by using quadratic, e.g. nonlinear combinations of the starting fermionic operators in defining new non-fermionic operators Âi . | cond-mat.str-el |
In this case the Kohn-Sham-Dirac equations ( 1) can be conveniently solved by expanding the spinors Φ λ (r) in a plane-wave basis. We discretize real space: r → r ij = (x i , y j ), x i = iδx, y j = jδy with i = 1...N x and j = 1...N y . Here δx × N x = δy × N y = L. Fourier transforms f (k) of real-space functions f (r) are calculated by means of a standard fast-Fourier-transform algorithm 18 that allows us to compute f on the set of discrete wavevectors k ij , k ij = (k x,i , k y,j ) = 2π L (n x,i , n y,j ) ,(22) with -N x /2 ≤ n x,i < N x /2 and -N y /2 ≤ n y,j < N y /2 (or, equivalently, 0 ≤ n x,i < N x and 0 ≤ n y,j < N y ), that belong to the Bravais lattice of the discretized box. The definition of the Fourier transform that we use is the following: f (r) = d 2 k (2π) 2 f (k) e ik•r f (k) = d 2 r f (r) e -ik•r . (23 ) After discretization f (r) → f ij = f (r ij ), f ij = f (k ij ) with f ij = 1 L 2 Nx-1 n=0 Ny-1 m=0 f nm e iknm•rij(24) and In all the numerical calculations reported on below we use L as unit of length, 2π /L as the unit of momentum, and v/L as the unit of energy. In what follows we also set = 1. f ij = L 2 × 1 N x N y Nx-1 n=0 Ny-1 m=0 f nm e -ikij •rnm . | cond-mat.str-el |
32 In the isotropic lattice magnetic ordering is suppressed and the reentrant behavior disappears. To illustrate the first-order nature of the metalinsulator transition we show in the upper panel of Fig. 3 the spectral weights of the cluster sites at E F = 0 as functions of U . The lower panel shows the average double occupancy d occ = i n i↑ n i↓ /3. Both quantities exhibit hysteresis for increasing and decreasing U , indicating co- energies below and above the Mott transition for T = 0.02t and t ′ = 0.8t. Plotted is the average over the three inequivalent sites within the unit cell. | cond-mat.str-el |
In topological systems like FQH liquids, edge states have been demonstrated to be very effective and essential probes of the bulk topological order in both theoretical calculations and experiments. As shown in Fig. 1, one expects two counter-propagating edge modes originating from the two edges at 2/3 → 1 and 1 → 0. 11,12,13 In the same spirit as the analysis in Refs. 33 and 34, we can label each low-energy edge excitation by two sets of occupation numbers {n L (l L )} and {n R (l R )} for the inner and outer edge modes with angular momenta l L , l R and energies ǫ L , ǫ R , respectively. n L (l L ) and n R (l R ) are nonnegative integers. | cond-mat.mes-hall |
Using Feynman's rules and the correspondence above, it is possible to establish the next equation for the diagrams shown in Fig. 1: = στ σ ′ τ ′ G 0 G στ σ ′ τ ′ + στ σ ′ τ ′ G 0 G 0 g 0 λ λ 1 2 + στ + σ ′ τ ′ g 0 g 0 λ λ λ λ G 0 G 0 G 0 1 2 3 4 - G 0 στ σ ′ τ ′ στ σ ′ τ ′ 4 1 G 0 g (0)ir 2 [1, 2|3, 4] g (0)ir 3 [1, 2, 3|4, 5, 6] 1 3 6 4 G 0 G 0 + -1 -1 2 G 0 G 0 λ 4 3 2 G 0 λ 6 2 5... = στ σ ′ τ ′ g 0 g στ σ ′ τ ′ + στ σ ′ τ ′ g 0 g 0 G 0 λ λ 1 2 + στ + σ ′ τ ′ G 0 G 0 λ λ λ λ g 0 g 0 g 0 1 2 3 4 - G 0 2 4 6 σ ′ τ ′ λ 3 5 G 0 G 0 3 g (0)ir 2 [4, 5|6, σ ′ τ ′ ] + στ +1 λ 3 1 g (0)ir 2 [στ, 1|2, 3] G 0 σ ′ τ ′ 2 στ -1 λ 2 1 g (0)ir 2 [στ, 1|2, σ ′ τ ′ ] G 0 + στ σ ′ τ ′ - 1 2 λ 4 1 2 4 3 G 0 ... g (0)ir 3 [στ, 1, 2|3, 4, σ ′ τ ′ ] FIG. 2: The diagrams for impurity electron propagator g σσ ′ (ττ ′ ). The last three diagrams contain the correlation contributions. Two of them are strong connected and the last is weak connected. G σσ ′ (τ -τ ′ |λ) = G 0 σσ ′ (τ -τ ′ ) + σ1σ2 β 0 dτ 1 β 0 dτ 2 G 0 σσ1 (τ -τ 1 )λg σ1σ2 (τ 1 -τ 2 |λ)λG 0 σ2 σ ′ (τ 2 -τ ′ ). | cond-mat.str-el |
The effect of antisymmetric DM and symmetric KSEWA interactions in spin ladders in the weak-coupling limit has been studied theoretically in Ref. 15. It was shown that in spin systems with SU(2) symmetry the DM term alone breaks this symmetry, opening a gap in the excitation spectrum above B c1 . On the other hand, the effect of KSEWA interactions is to recover the SU(2) symmetry, leaving the excitation spectrum incommensurate but gapless. No sign of a gap in the excitation spectrum has been detected by NMR 39 in the vicinity of B c1 and B c2 down to 40 mK confirming the applicability of the LL formalism for describing the intermediate phase. The presence of KSEWA interaction on the ladder rungs or a combination of the DM and KSEWA interactions on the ladder legs (resulting in the anisotropy observed by ESR, but with a gapless excitation spectrum in the fieldinduced intermediate phase) explains the experimental results. | cond-mat.str-el |
However, for the cluster, due to the imposed boundary conditions, the case U=0 describes a well localized state of the hole; in this sense it is bound. Our numerical results qualitatively agree with those of previous publications 1,2,3,4 . A detailed quantita-tive comparison is not possible because the previous publications have considered spin symmetric clusters while we impose a spontaneous violation of the SU(2) symmetry via boundary conditions. The dispersion of a free hole for various values of t and t ′ is well known from previous work 12,13,14 . The bound state results in Table I for values of t and t ′ that correspond to the free hole dispersion minima at the nodal points, (±π/2, ±π/2), are presented by large font. For all other values of t and t ′ the dispersion minima are at the antinodal points, (±π, 0), (0, ±π), or at the Γ-point, k = 0. | cond-mat.str-el |
Figure 1 shows the cantilever sensitivity T/To as a function of the local slope ρ rms . In fact, Eq. ( 5) defines a limiting value of the local slope ρ rms for which T=0, yielding 111)) 30 and v L =0.28 (Si( 100)) 30 we obtain respectively ρ rms/max =2.13 and ρ rms/max =1.6. For a metallic overlayer as gold (widely used to coat cantilevers) with v L =0.44 30 we obtain ρ rms/max =1.12. These are relatively significant values for ρ rms and the pertubative expansion of Eq. ( 5) are valid only for local slopes ρ rms < 1. | cond-mat.mes-hall |
Instead of 12 level crossings, there are only six, each corresponding to a jump ∆S z = 2 occurring at fields gµ B H (k) flip = 1 N k (E 0 (S z ) -E 0 (S z -k)) , k = 2. (10) The saturation field is now given by an instability criterion of the fully polarized state towards a two-magnon excitation, gµ B H ck = 1 N k E 0 N 2 -E 0 N 2 -k , k = 2. (11) For J 1 < 0, this field is larger than the field of the onemagnon instability given by the above equation with k = 1 and therefore determines the predominant instability when lowering the field in the fully polarized state. A necessary condition for a ∆S z = 1 level crossing to occur is that the lower bound E 0 (S z ) of the energy spectrum at zero field for a fixed value S z is a convex function of S z , i. e., the condition E 0 (S z + 1) ≤ 1 2 (E 0 (S z ) + E 0 (S z + 2))(12) must be fulfilled at H = 0. At the special point J 1 = 0, J 2 > 0 (φ = π/4), the J 1 -J 2 lattice decouples into two independent Néel sublattices, and equality holds above. For the finite size clusters which we consider, enlarging φ further (i. e. making J 1 ferromagnetic) stabilizes two-magnon bound states, and Eq. | cond-mat.str-el |
This may be related to the fact that dot pattern production by IBS is relatively recent. In particular, the dependence of the morphological properties of these patterns with temperature, ion energy and fluence has to be systematically addressed in order to obtain a more general picture of the nanopatterning process. On the other hand, both types of IBS patterns share many behaviors, which is consistent with the corresponding theoretical models. Also, there is a further experimental issue to be addressed: the possible influence of technical parameters on IBS pattern production. This important problem has been only studied by Ziberi and coworkers [194]. They obtained that the settings of the Kaufman ion-gun, the one that they employed, can affect the IBS nanopattern. | cond-mat.mtrl-sci |
In the weak coupling regime (Fig. 3a and Fig. 3b 1 ) POC at frequencies around 2J is shifted down by about 1% indicating the clear magnetic origin of this peak in the weak coupling regime. To the contrary, POC LOW , with frequency around ω 0 , is shifted down about 6% (Fig. 3b 2 ), that is just softening of ω 0 induced by IS. Thus, we get one more confirmation of the phononic origin of theoretical POC LOW and experimental MIR LOW band, in agreement with Ref. | cond-mat.str-el |
First, we observe that the spin correlation in the z-direction shows exponential decay with correlation length which does not change with z. Second, the spin correlation shows little size dependence in the x and y direction. Therefore, from the results of the short-range spin correlation and the scaling relation as shown in Fig. (1a), we concluded that our calculations already represent the thermodynamic limit. ξ(T, K) is computed from ξ(L, T, K) by taking the mean of the values for L = 4 and L = 5. We plot ξ(T, K) as a function of T for those field values in Fig. | cond-mat.str-el |
In particular, we assume that they traverse slowly a closed contour in the X -Z plane, as illustrated in Fig. 4. Then the pseudo-spin orbital eigenstates states follow their respective pseudo-fields. In addition to the dynamical phase each state acquires a Berry phase. As one can see from Fig. 4, the Berry phases acquired by the two ground states from the subspaces "+" and "-" are opposite in signΦ ± = ± 1 2 dϕ cos θ ,(11) where the angles ϕ and θ are introduced in Figs. | cond-mat.mes-hall |
The CDW2 region now appears at positive U and V , contrary to the results in the previous figures, as long as the stabilizing interaction V is large enough. The boundary that separates CDW2 from PM in Fig. 10 is very close to QMC results 23 , which gives us confidence in the validity of the results. The ferromagnetic phase is dominant when both U and V are large. The ICDW phase does not appear at filling n = 2/3 (Fig. 11). | cond-mat.str-el |
The spatial distributions of the force stresses near the dislocation line are presented in figure 3. The force stress fields have no artificial singularities at the core and the maximum stress occurs at a short distance away from the dislocation line (see figures 4 and 5). They are zero at r = 0. It can be seen that the stresses have the following extreme values: |σ xx (0, y)| ≃ 0.546A/ℓ 2 + 0.260B/ℓ 3 at |y| ≃ (0.996ℓ 2 + 1.494ℓ 3 )/2, |σ yy (0, y)| ≃ 0.260A/ℓ 2 -0.260B/ℓ 3 at |y| ≃ (1.494ℓ 2 + 1.494ℓ 3 )/2, |σ xy (x, 0)| ≃ 0.260A/ℓ 2 + 0.546B/ℓ 3 at |x| ≃ (1.494ℓ 2 + 0.996ℓ 3 )/2, |σ yx (x, 0)| ≃ 0.260A/ℓ 2 -0.260B/ℓ 3 at |x| ≃ (1.494ℓ 2 + 1.494ℓ 3 )/2 , and |σ zz (0, y)| ≃ 0.399A at |y| ≃ 1.114ℓ 2 . Thus, the characteristic internal lengths ℓ 2 and ℓ 3 determine the position and the magnitude of the stress maxima. For γ > 0 the stresses σ xx and σ xy are bigger than in the case γ = 0 and σ yy and σ yx are smaller than in the case γ = 0 (see figure 5). | cond-mat.mtrl-sci |
In the cup-shaped potential U e the electrons will be found with the greatest probability where the potential has a minimum. This means that the extra charge on the helicoid will concentrate in a strip around the value of ξ min , i.e. a solution to dU e /dξ = -eE. n = 1 n = 2 n = 3 n = 10 m = 0 0. Application of an electric or magnetic field along the x-axis would nontrivially affect the motion of electrons on the surface-this problem will need to be studied numerically. It would be very interesting to observe the predicted effect in graphene ribbons or helicoidal ribbons synthesized from a semiconducting material. | cond-mat.mes-hall |
For the current site, the total symmetry Q tot of the full quantum many-body state manifests itself in that the corresponding A-matrix fulfillsQ l + Q r + Q σ = Q tot . For the handling of matrix product states quantum numbers imply a significant amount of bookkeeping, i.e. for every coefficient block we have to store its quantum number. The benefit is that we can deal with large effective state spaces at reasonable numerical cost. The Lanczos algorithm, in particular, takes advantage of the block structure. Of course, the treatment of Abelian symmetries is generic and not limited to only one symmetry. | cond-mat.str-el |
Hence a candidate RDD variable (denoted as κ4 ) which provides cancellation of singular terms in the Pauli potential could be defined as κ4 = s 4 + b 2 p 2 , b 2 > 0 . (61) Note that this form is manifestly positive. This RDD can be used to construct a variety of enhancement factors to replace Eq. ( 45) for the fourthorder approximation to the Pauli term, for example F θ (κ 4 ) = a 4 κ4 . This simplest enhancement factor corresponds to a Pauli potential with finite values at point nuclei but clearly it is not the only κ4 -dependent one with that property. Any linear combination of non-singular enhancement factors (including the simple F θ = 1) also will be non-singular. | cond-mat.mtrl-sci |
Quantum number fractionalization is one of the most intriguing topics in physics. In particular, searching for the magnetic monopole has been of persistent interest for centuries. This interest was recently aroused again by the discovery of the magneticmonopole-like spin excitations in the spin ice systems [1,2]. The energy of the monopole-anti-monopole pair was shown to be inversely proportional to their distance of separation, i.e. the Coulomb phase of the magnetic monopole [3]. The spin ice systems are found in the pyrochlore lattice. | cond-mat.str-el |
(ii) The resulting equations are Volterra type integro-differential equations, for which highly stable and accurate algorithms can be found in the literature, 46 and which remain causal even when they are approximated numerically. Finally, (iii), the real-frequency representation which we introduce in Sec. IV D to handle initial states at zero temperature is based on this approach. In the following we assume that Y and K satisfy the hermitian symmetry (10), such that it is sufficient to determine the Matsubara, retarded, mixed "¬", and lesser components of Y . Corresponding components of the convolution K * Y in Eq. ( 37) are obtained from the Langreth rules, 35 which follow directly from the definitions (3)-( 8) and the definition of the contour integral. | cond-mat.str-el |
The maximum thickness of VA-CNTs was 1.2 mm at a Fe thickness of 0.5 nm. The thickness of VA-CNTs decreased when Fe thickness exceeded 0.5 nm. Figure 1b shows TEM images of CNTs grown under the same condition as Fig. 1a on substrates with uniform Fe thicknesses of 0.5 and 1.0 nm. SWNTs with a diameter around 4 nm mainly grew for 0.5-nm-thick Fe catalyst, whereas thicker CNTs grew for 1.0-nm-thick Fe catalyst. This difference in CNTs is because a thicker initial Fe layer yields larger Fe nanoparticles, 8 indicating a narrow VA-SWNTs growth window for the initial Fe thickness. | cond-mat.mtrl-sci |
Although Fe and Nb are both type (iii) elements, the experimental observations can be explained by the fact that Fe and Nb actually behave like a type (ii) and type (i) element at high temperatures, respectively (see Fig. 2). Finally, Fig. 4 shows our predicted site occupancy of Ta, Zr, Nb, Mo, V, Cr, Mn, Fe and Ni in Ti 0.47 Al 0.51 X 0.02 alloys in direct comparison with the experimental data from Hao et al. [27]. All calculations are performed at the experimental annealing temperature of 1173K. | cond-mat.mtrl-sci |
( 48) and is resposible for the identical spin orientation for two spin splitting hole states in small k regime. The spin splitting between HH± depends on the coupling between |1, 3 2 and |1, -3 2 through higher-order perturbation. Different from the electron case, the direct coupling will not cause the x-direction or y-direction spin orientation. Instead, it results from the coupling between |1, 3 2 (|1, -3 2 ) and |1, 1 2 (|1, -1 2 ). For two LH1 states, denoted as Ψ lh,± , such coupling will lead to the spin orientations of Ψ lh,+ opposite to Ψ hh,+ , and that of Ψ lh,-opposite to Ψ hh,-. Thus the total spin orientation of the 2DHG is conserved, though Ψ hh,+ and Ψ hh,-have the same spin orientation in the low hole density regime. | cond-mat.mes-hall |
On the BCS side the interaction parameter U is negative (the atomic interaction is attractive). For simplicity we assume that each well of the periodic potential for atomic motion in three dimensions could be approximated by a harmonic potential. This harmonic approximation gives the following analytical results of J and U ( = 1) 6 : J = E R e -π 2 √ s 4 π 2 s 4 - √ s 2 - s 2 1 + e - √ s , U = - 8 √ π |a s | λ 2s 3 E 3 R mλ 2 1/4 . Here λ is the laser wavelength, s is the lattice height, and m is the mass of the trapped 6 Li atoms. The recoil energy of the lattice E R = π 2 /2md 2 depends on the lattice constant d = λ/2. In our numerical calculations the wavelength is chosen to be λ = 1030 nm (E R = 1.293 × 10 -11 eV) 6 . | cond-mat.str-el |
Assuming that the device width is large and homogeneous along the direction transverse to current flow, we can write the Hamiltonian as H = 2 6 6 6 6 4 α β1 β † 1 α β2 β † 2 α . . . . . . | cond-mat.mes-hall |
For instance, by defining H ± = e i √ 8 ϕ e ± i 2 φ e -i √ 3 8 χ1∓ i 2 χ2 ,(54) the holes have fermionic monodromies and their OPE's are given by H ± (z)H ± (w) ∼ (z -w) H -(z)H + (w) ∼ (z -w) 0 ,(55) where only the leading short distance behavior is indicated. We could equally well have chosen a bosonic representation of the quasiholes, in which case the operators are given by H ± = e i √ 8 ϕ e ± i 2 φ e -i √ 7 8 χ1∓i √ 3 2 χ2 ,(56)and the OPE's by H ± (z)H ± (w) ∼ (z -w) 2 H -(z)H + (w) ∼ (z -w) 0 . (57) Using these hole operators, the monodromies will be either absent or just a sign, corresponding to bosons and fermions respectively (and in the latter case it is also important to remember that the holes are not allowed to be at the same point). In the Ising representation, on the other hand, the hole wave functions have non-Abelian monodromies. Thus there must be a corresponding non-Abelian Berry matrix in the 331 representation. In the following, we will use the fermionic representation of the quasiholes. | cond-mat.mes-hall |
We note that with the increase of temperature, the Rabi frequency drops with the decline of the amplitude. Hence we plot in Fig. 3 the average occupation of state |i [1/t ∞ t∞ 0 P i (t)dt] (i = 0, 1, 2) as function of the temperature, which is obviously show that it is less possible to create the state with one electron in the second dot (state |2 ) when the temperature is high. On the other hand, the environment temperature will also affect the beat pattern, which is caused by the electronphonon coupling. with the rise of the temperature, the beat pattern will decay and the population of the state |2 will decrease as well. As the temperature reachs about 50ω 0 , the beat pattern almost disappears. | cond-mat.mes-hall |
As already mentioned above it is quite interesting that the appearance of an additional dilatational strain leads to an effective anisotropy of the system: Whereas in the case of a positive admixture ǫ + the velocity grows with increasing η, the opposite happens for negative admixture ǫ -. This suggests that in free growth the bicrystal structure will eventually grow in the direction with the higher propagation velocity. For the fast branch, we did not find solutions FIG. 9: The left panel shows the stability parameter σ * as function of the control parameter ∆ el /p, for positive mixing, whereas the in the right panel σ * (∆ el /p) is plotted for negative mixing. Additionally, both graphs contain the σ * (∆ el /p) dependence for the pure shear configuration η = 0, as a reference to compare both types of mixings, which was taken from [10].beyond the point η ≈ 0.25. In Fig. | cond-mat.mtrl-sci |
( 4)-( 5), we have derived the following expression for the electron self-energy in the chiral basis due to the phonon-mediated e-e interaction Eq. ( 5): Σk± = -k B T g 2 2 λ q,iqn D 0 (iq n ) G0 k+qλ (ik n + iq n ) [1 ∓ λcos(φ k+q -2φ q )] ,(6) where G0 kλ (ik n ) = 1/(ik n -ξ kλ ) is the non-interacting electron Green function in the chiral basis, with ξ kλ = λǫ k -ε F the quasiparticle energy rendered from the Fermi level. The off-diagonal elements of the self-energy matrix Σk±∓ , which couples the conduction band and valence band, are found to be zero after performing the angular integration (this is true also for the pure Coulomb interaction case [5]). For TO phonons, Eq. ( 6) remains the same except the angular factor becomes 1∓cos(φ k+q -2φ q ) → 1±cos(φ k+q -2φ q ). The Matsubara sum in Eq. | cond-mat.mtrl-sci |
1 Up to now implementations of the LDA+DMFT approach utilized linearized and higher-order muffintin orbital [L(N)MTO] techniques 17 and focused on the investigation of electronic correlation effects for a given lattice structure. However, the mutual interaction between electrons and ions, i.e., the influence of the electrons on the lattice structure, is then completely neglected. LDA+DMFT computations of the volume collapse in paramagnetic Ce 18,19 and Pu 20 and of the collapse of the magnetic moment in MnO 21 did include the lattice, but only calculated the total energy of the correlated material as a function of the unit cell volume. 22 In the case of more subtle structural transformations, e.g., involving the cooperative Jahn-Teller (JT) effect, 23,24 the L(N)MTO technique is not suitable since it cannot reliably determine atomic positions. This is due to the atomic-sphere approximation within the L(N)MTO scheme, where a spherical potential inside the atomic sphere is employed. Thereby multipole contributions to the electrostatic energy due to the distorted charge density distribution around the atom are ignored. | cond-mat.str-el |
The observation of scale invariance suggested that fluctuations should not be imputed to the discrete nature of the deforming lattices. Instead, the collective behavior of crystal dislocation was held responsible for the emergence of scale-free patterns, in a close-to-critical behavior fashion. Such an observation triggered a number of numerical [2,9,10] and, whenever feasible, analytical [11,12] studies so that the validity and the limits of applicability of this picture could be assessed. At present, there is evidence that the notion of scale invariance in crystal plasticity as a consequence of a close-to-critical state is consistent with observation in single and multi-slip geometries, over a variety of materials, hardening coefficients and loading conditions. However, the general understanding of these processes is still chiefly phenomenological and several aspects of such phenomenology are still obscure. In particular, it is still not entirely clear to which extent such considerations should be extended to the case of polycrystalline geometries. | cond-mat.mtrl-sci |
The only example of such a state has been found in the model which physical parameters are dramatically different from those for the cuprates [25]. Ironically, this is a situation typical for the cuprate physics: those models which allow rigorous treatment do not apply and those models which may apply do not allow rigorous treatment. The other difference comes from the influence of gapless excitations. In [16] the interaction of quasiparticles with gapless collective modes strongly renormalizes the self energy resulting in a marginal Fermi liquid state. In [23] the interaction between gapless gauge field excitations and quasiparticles is weak and the state is essentially a Fermi liquid. II. | cond-mat.str-el |
Introduction Very peculiar magnetic, thermal, and transport properties of 4f electron based heavy fermion systems are determined by the interplay of the strong repulsion of 4f electrons on the rare-earth ion sites, their hybridization with wide band conduction electrons, and an influence of the crystalline electrical field. Consequences of the mentioned interplay for the electronic energy band structure near the Fermi-energy (E F ) were recently studied in YbRh 2 Si 2 and YbIr 2 Si 2 by angle-resolved photoemission and interpreted within the periodic Anderson model [1,2]. It was found that the hybridization of 4f electrons results in a rather flat 4f band near E F . Additionally, renormalization of the valence state leads to the formation of a heavy band that reveals strong 4f character close to E F . Moreover, slow valence fluctuations of the Yb ion may occur between 4f 13 and closed 4f 14 configurations with an averaged valence value of about +2.9 [3]. Evidently, these observations are consistent with a metallic behavior with very heavy charge carriers having properties of a Landau Fermi-liquid (LFL). | cond-mat.str-el |
The different parameter windows for pattern formation, smoothening or roughening have been studied by Chason et al. [31] for Cu(001) surfaces as a function of ion dose and target temperature. In this scheme, the surface remains flat for low ion fluxes and high temperatures whereas it roughens for high fluxes and low temperatures. For intermediate temperatures (200-350 K) the diffusive regime dominates and, finally, high ion fluxes and high temperatures imply an erosive regime. Pattern formation in thin metal films by ion beam sputtering In the previous section, we have seen that ripples can be also produced by IBS on single-crystal metals for which ES energy barriers [54,150] play such a crucial role, influencing the stability of the ripple morphology even at room temperature [177]. However, a different scenario has to be considered for IBS nanostructuring of thin metallic films. | cond-mat.mtrl-sci |
Nowadays, with the great improvement of energy resolution, ARPES is able to observe the fine structures due to e-ph interaction 2 and the tiny superconducting gap 3 with an energy resolution less than 1 meV. Recently, the oxygen isotope effect has been studied with ARPES on the high-T c superconductor Bi 2 Sr 2 CaCu 2 O 8 (Bi2212) by two groups 4,5 . Comparing their data, at least one common feature has become clear, the spectra are shifted when 16 O is substituted by its isotope 18 O, whose movement in this crystal is regarded as a lattice vibration, i.e., phonon. This isotope induced band shift indicates that the e-ph interaction has an effect on the electrons, hence providing direct evidence for the interplay between electrons and phonons in this material. Since the first report by Gweon et al. 4 , the isotopic band shift has become an controversial issue and gained considerable interest 7,8,9 , as the observed band shift is up to 40 meV, much larger than the maximum isotopic energy change of phonon (∼ 5 meV) according to the measured vibration energies of oxygen 6 . | cond-mat.str-el |
In Fig. 3 (from Ref. [33]), the main peak of the polaron optical absorption for α = 5 at Ω = 3.51ω LO is interpreted as due to transitions to a RES. The "shoulder" at the low-frequency side of the main peak is attributed to mainly one-phonon transitions to polaron-"scattering states". The broad structure centered at about Ω = 6.3ω LO is interpreted as a FC band. As seen from Fig. | cond-mat.mes-hall |
For the sake of simplifying the notation, we will employ in this section a 2 × 2 matrix representation which is detailed in the Appendix. The correlations between the angular deviations δφ and δθ are well described by the lesser magnon Green function D < (r, t, r , t ) defined as i D < = δθ(r , t )δθ(r, t) δφ(r , t )δθ(r, t) δθ(r , t )δφ(r, t) δφ(r , t )δφ(r, t) . The representation of the lesser function in the momentum space D < (k, t, k , t ) is obtained by expanding δθ and δφ on the wavefunctions ξ k , according to ( 18) and ( 19). The representation in momentum space is quite useful as it reveals the relationship between the angular deviations δφ,δθ and the statistics of the magnons. The force F H depends on the diagonal component of the lesser function D < = D < θθ = D < φφ , F H = kq f H (k, q)i D < (k -, t, k + , t) ,(80) where k -= kq/2 and k + q/2 and f H (k, q) = - K u 3λ 2π L 1 i sech πλq x 2 λq x (81) × 1 + 3(k x λ) 2 + (qxλ) 2 4 √ ω k-ω k+ ν θ k-ν θ k+ + ν φ k-ν φ k+ . The force F j depends on the off-diagonal component of the lesser function D < off = D < φθ , F j = kq f j (k, q) d dt i D < off (k -, t, k + , t),(82)withf j (k, q) = -Sv -k-,k+ . | cond-mat.mtrl-sci |
An applied side gate voltage increases this value, until it almost reaches 8e 2 /h in the case of V SG = ±60 V. This shows that it is possible to increase the minimum value of conductivity by creating a density gradient along the width of the conductor. C. Side gates in p-n configuration Besides having all side gates at the same potential, many other, more complicated configurations are possible with four side gate electrodes. Here we focus on the situation where the in-plane gates are biased such that a p-n-like configuration along the Hall bar arises. The same voltage is applied to side gates on opposite sides of the Hall bar. For both pairs the applied voltage has the same absolute value, but opposite sign [Fig. 3(a)]. | cond-mat.mes-hall |
For the fibre and ITZ, the thermal expansion coefficient and Young's modulus in Tab. 2 for T = 20 • C was assumed to be independent of the temperature. RESULTS The results are divided into two parts. Firstly, the effect of radial stresses are investigated. Secondly, the influence of the fracture process zone is studied. Cooling down of the droplet is expected to generate large radial stresses. | cond-mat.mtrl-sci |
The normal-electronic nematic transition The nematic order parameter in 2D is a l = 2 representation of the SO(2) rotational group [33]. It is defined as a symmetric traceless tensor of rank two. N = n 11 n 12 n 12 -n 11 . (4.1) The 2D rotational group SO(2) is isomorphic to U (1). Hence, we define instead the complex order-parameter field N ( r, t) N ( r, t) = n 11 ( r, t) + in 12 ( r, t),(4.2) where r and t are the space and time coordinates. We will use this complex order parameter field in this paper to take the advantage of the Abelian nature of SO (2). | cond-mat.str-el |
1, bottom). Metal surfaces behave in the same way, although the electrons available for donation into oxidising molecules are rather uniformly delocalised over the atoms composing the surface. The reactions between O 2 molecules and bare metal surfaces have been studied using scanning tunneling microscopy (STM) under ultrahigh vacuum conditions [38][39][40]. When a previously cleaned metal surface is treated with a flow of oxygen gas at low pressure, pairs of oxygen atoms separated by distances of about 1 to 3 nearestneighbour metal-metal distances become visible under the STM (Fig. 2). This suggests that O 2 reacts dissociatively with metal surfaces by cleavage of the O-O bond, leading to pairs of single adsorbed oxygen atoms. | cond-mat.mtrl-sci |
The advantage of graphene with spin gap asymmetry, i.e. different ∆ α and ∆ β gaps, found in this work is the different values of the critical electric field required to close these gaps, such that | E c(β) |<| E c(α) | when ∆ β < ∆ α . Therefore, this structure will be characterized by the spin-polarized current and by a non-symmetric current-voltage characteristics as for a semiconductor diode, when the current flow in one direction is preferable to the other. IV. DOPING OF GRAPHENE We have also investigated the influence of impurities on the electronic structure of graphene in the case when they are not embedded at the zigzag edges. Replacing carbon atoms by nitrogen atoms in a graphene lattice results in the appearance of impurity levels inside of both the ∆ α and ∆ β gaps. | cond-mat.mtrl-sci |
Next order contribution ρ(-1) satisfies the equation I (2) col (ρ (-1 neq ) + I (3) col (ρ (-2) neq ) = 0, (37) where ρ(-2) is already found by solving (36). It turns out that ρ(-1) neq , found from (37), is still diagonal in band indexes and contains the antisymmetric contribution in the transverse to the electric field direction. It leads to the transverse conductivity that, like ρ(-2) neq , depends as 1/n on the impurity concentration n. At zeroth (next) order in the disorder strength, both inter-band and intra-band matrix elements become important. One can separate 4 distinct parts. ρ(0) neq = ρint + ρsj + ρadist + ρsk ,(38) where first two terms are purely off-diagonal and the other two are diagonal in band indexes. These contributions to the density matrix satisfy following equations. | cond-mat.mes-hall |
We have shown simulated curves with second order model only, because the first order model failed to converge at low fields values and is largely off from the measurements. Apart from a constant serial resistance (0.5 Ω) than higher systematically the experimental bias, the model was able to reproduce nicely the experiment, in particular for high temperature where the transport in the structure is clearly dominated by optical phonons. The model predicts the low temperature curve with less accuracy because the transport at such a temperature also require the computation of scattering rates due to acoustical phonons and electron-electron interactions, which are not computed in the present model. Agreement between computed and experimental current-voltage characteristics has already been reported for other model approachs such as based on Monte Carlo 13 or non-equilibrium Green's functions 14 ; the comparison was done however on a much more limited range of currents, temperature and structure design. Formally, the current driven by tunneling between two subbands through a barrier or by optical absorption are physically equivalent because both processes conserve the in-plane wavevector. As a result, the striking agreement between the predictions of the second-order model and the experiment can be interpreted as a strong experimental evidence for the validity of the Bloch gain model. | cond-mat.mes-hall |
Dirac) point. In particular, early theoretical work 28,37,38,39,40 predicted a universal T = 0 minimum conductivity σ min = 4e 2 /πh at the graphene Dirac point in clean disorder-free systems. The inclusion of disorder-induced quantum anti-localization effect, assuming no intervalley scattering, leads to a theoretical infinite minimum conductivity at the Dirac point, whereas the presence of inter-valley scattering localizes the system leading to zero conductivity at the Dirac point. This confusing theoretical picture stands in stark contrast to the experimental reality, where the graphene conductivity is approximately a constant (as a function of gate voltage or carrier density) around the Dirac point, with this constant minimum conductivity plateau having a nonuniversal sample dependent value (∼ 4e 2 /h -20e 2 /h). It was first suggested in Ref. 22 that the minimum conductivity phenomenon is closely related to the break-up of the graphene landscape into inhomogeneous puddles of electrons and holes around the Dirac point due to the effect of the charged impurities in the environment. | cond-mat.mes-hall |
1(d-f) show an enlarged view of the 1st ML of growth from Figs. 1(a-c). At 1000 • C, diffuse scattering appears between the first and second pulses. At sufficiently lower temperatures (≤ 785 • C), diffuse scattering is not visible until after the second pulse, indicating either delayed nucleation or intensity below our detection limit. To extract quantitative information, the x-ray data were fit to the sum of three independent components, I f it (q || ) = I bg + I spec (q || ) + I dif f (q || ± q 0 ). (1) In this equation, I bg is a constant background, and I spec and I dif f take the form: f (x) = I 0 /[1 + ξ 2 x 2 ] 3/2 , (2) where ξ is the correlation length. | cond-mat.mtrl-sci |
These authors were the first to clearly identify sharp finite voltage bias features as a Kondo effect and not as simple cotunneling via excited states. The main idea behind Kondo physics is the existence of a degeneracy, which is lifted by the conduction electrons. This is clearly the case for a quantum dot with only one electron on the last orbital, leading to a doubly degenerate spin S = 1/2. For a quantum dot with two electrons and two nearly degenerate orbital levels, two different kinds of magnetic states occur: a singlet and a triplet. Depending on δE the energy difference between the two orbital levels and J the strength of the ferromagnetic coupling between the two electrons, the splitting between the triplet and the singlet can in principle be tuned, and eventually brought to zero, leading to the so-called singlet-triplet Kondo effect [25]. However the singlet is in most situations the ground state, leaving the triplet in an excited state, thus suppressing the Kondo effect. | cond-mat.mes-hall |
The answer is no, not necessarily: in Sec. II we provide an example, the 1/r fermionic Hubbard model [22], which shows that relaxation in the presence of a Mott gap is indeed possible. Note that the formation of a fermionic Mott insulator was recently observed with ultracold atoms [23]. Another central question is whether the steady state of a quenched isolated system can be described by an effective density matrix ρ, such that Tr[Oρ] yields the correct expectation value for any observable O which relaxes. Statistical mechanics can be used to make an approxi-mate but usually accurate prediction ρ mic for this steadystate density matrix. For example, the microcanonical prediction is that ρ mic = const for states with energy close to Ψ(0)|H|Ψ(0) , and zero otherwise. | cond-mat.str-el |
However, this often comes at a cost of reduced accuracy, and so data is generally collected for at least one applied field using the standard relaxation technique to compare the two methods directly. In the present case, this was done at 33 T (black triangles in Fig. 2(b)). The excellent agreement between the techniques suggests that the dual slope method estimates the heat capacity well in Sr 3 Cr 2 O 8 .∆S = κ(T -T bath ) T dt(2) The emergence of a lambda-like anomaly was clearly observed with an increasing applied field beyond H c1 ∼ 30.4 T; this provides unambiguous evidence for the field-induced order in this system. The lambda anomaly is small for fields close to H c1 , but becomes much more prominent with increasing applied field as seen in Fig. 2(b). | cond-mat.str-el |
Consider a four-fermion interaction with incoming momenta K 1 and K 2 , and outgoing momenta K 3 and K 4 . The difference between incoming and outgoing momenta at the left vertex can be small, say K 3 -K 1 ≡ q left vertex ≈ 0. This can match up with small momentum transfer on the right: K 4 -K 2 ≡ q right vertex ≈ 0. However, this says nothing about the relationship between K 1 and K 2 . Indeed, K 1 and K 2 can each independently take any value around the Fermi surface, i.e. | K 2 -K 1 | can take any value between 0 and 2K F . | cond-mat.str-el |
2 (For λ = 1, there is a second bound state between the one shown and the continuum, therefore the bandwidth is much less than Ω). For MA (1) and MA (2) this problem is indeed fixed, and the polaron dispersion width (at weak couplings) is Ω. All other quantities are also clearly more accurate. C. Polaron+one-phonon continuum and higher energy states In order to understand the effects on higher-energy states, we study the spectral weight A(k, ω) = -1 π ImG(k, ω). As is well known, this is finite only at energies ω where eigenstates of momentum k exist. For discrete (bound) states the spectral weight is a Lorentzian of width η and height proportional to the qp weight. | cond-mat.str-el |
Finally, the interaction between the pristine 10-AGNR and the gas molecules is written as a sum over each molecule:V = ℓ vℓ ,(5)with vℓ = - Ns i=1 Nm n=1 t in (ĉ † i dn + d † n ĉi )δ i,i ℓ δ n,n ℓ ,(6) where the indice n ℓ labels the atomic site n of the molecule ℓ that is adsorbed on site i = i ℓ of the ribbon, and t in = τ is the hopping integral between the ribbon and the molecule. N m is the number of atoms at each molecule. Now we can calculate the Green's function of the whole system to obtain the average LDOS. For this purpose, as mentioned above, we divide the system into an infinite series of supercells. The supercells are labeled by ℓ = 0, ±1, ±2, • • • and we focus on one of them, say ℓ=0, and consider the supercells ℓ = -1, -2, • • • and ℓ = +1, +2, • • • as neighboring supercells at the left and right, respectively. Based on the iterative procedure introduced by López Sancho et al. | cond-mat.mes-hall |
S str -this is shown in Fig. 12 for MgO (32 atoms/cell) and MgNH (12 atoms/cell). Like with the degree of order P, the correlation is worst for the AB 2 Lennard-Jones crystal, the reason being its instability to decomposition that leads to long-period layered structures that contain quite diverse atomic sites. The same can be expected for other complex and frustrated systems. A good example of which is the recently discovered [31] high-pressure stable phase of boron, the structure of which contains atomic sites that are so different that there is charge transfer between them [31] and the structure possesses large S str =0.18. which formulates an intuitively reasonable expectation that in stable structures atoms occupy. | cond-mat.mtrl-sci |
Both have an unambiguously s-shaped envelope function (not shown here). The splitting between both states (≈ 9 -10 meV) does not increase for smaller QDs, but is constant. It corresponds to the energy separation between A and B band in strained GaN. The higher excited hole states can not be unambiguously assigned to p-or d-like orbitals. Please note, that although they have been labeled according to the major band contributions, this contribution sometimes does not exceed 50 %. The QDs with aspect ratio 1:10 [Fig. | cond-mat.mtrl-sci |
The volume susceptibility of the washed cylindrical samples was lowest for the materials with the least amount of iron. The OFHC copper sample with 0.0002% Fe has a volume susceptibility of 3 × 10 -5 and Torlon 4301 with 0.0002% Fe has a volume susceptibility of 2 × 10 -5 . There is a discrepancy between the volume magnetic susceptibility of OFHC copper measured using the VSM and the textbook value of -1×10 -5 [17]. This could be due to an imperfect acid wash or the limited accuracy of the VSM of ±10 -5 for samples of this size. Despite this inaccuracy, the results of the VSM are show that an acid wash can reduce the volume suspectibility below 10 -4 which is sufficient for reducing heading error. Varying applied field measurements In addition to the low field volume susceptibility measurements, a complete scan of the magnetisation at different applied fields up to 1 T was preformed on the two 5 mm diameter Torlon cylinders. | cond-mat.mtrl-sci |
Within MA (2) , a true second bound-state is observed for the higher couplings shown in Fig. 9. In contrast to MA, which shows several bound states which disperse as k increases, MA (2) also clearly shows continua in between these discrete states. These account for some of the spectral weight that was in the MA peaks. These results again suggest a very fractionalized spectrum at intermediate and strong couplings. Instead of the polaron band, a second-bound state and a rather featureless continuum at higher energies, we instead find many sets of discrete states interspersed with continua. | cond-mat.str-el |
B † Q1 . . . B † QN |v v| B QN . . . | cond-mat.mes-hall |
I. INTRODUCTION Two-dimensional (2D), spin-1/2, Heisenberg antiferromagnets (HAFs) have been much studied in recent years. The interplay between (either dynamic or geometric) frustration and quantum fluctuations in determining the ground-state (gs) phase diagram of such models has been of particular interest. While such models are well understood in the absence of frustration, 1 this is not the case for frustrated systems, for which the zero-temperature (T = 0) phase transitions between magnetically ordered quasiclassical phases and novel (magnetically disordered) quantum paramagnetic phases 2,3 have become the subject of great recent interest. A particularly well studied such model is the frustrated J 1 -J 2 model on the square lattice with nearest-neighbor (NN) bonds (J 1 ) and next-nearest-neighbor (NNN) bonds (J 2 ), for which it is now well accepted that there exist two phases exhibiting magnetic long-range order (LRO) at small and at large values of α ≡ J 2 /J 1 respectively, separated by an intermediate quantum paramagnetic phase without magnetic LRO in the parameter regime α c 1 < α < α c 2 , where α c 1 ≈ 0.4 and α c 2 ≈ 0.6. For α < α c 1 the gs phase exhibits Néel magnetic LRO, whereas for α > α c 2 it exhibits collinear stripe LRO. | cond-mat.str-el |
In conclusion, we have studied the electronic structure of PCSMO (y = 0.25, 0.40) thin fillms by means of XPS. In the case of y = 0.40, the IMT is clearly observed on the Mn 2p core-level and valence-band spectra across the PI to FM transition. For y = 0.25, hysteresis is observed between 50 K and 125 K associated with the COOI-FM first-order phase transition. It is phase separated below 125 K and does not reach 100% FM even at the lowest temperature. Persistent PIPT from COOI to FM is also observed for y = 0.25 in the temperature range between 50 K and 70 K. After sufficient illumination, the system is as much metallic as allowed in the equilibrium, showing that the photons transform the bulk of COOI rather than connect the marginally disconnected metallic islands as is proposed for CMR1. The authors would like to acknowledge fruitful discussions with G. A. Sawatzky. | cond-mat.str-el |
Here,T K = D exp ǫ f N ρcV 2 is the single impurity Kondo temperature, where ρ c ≈ (2D) -1 is the density of states for conduction electrons with the half bandwidth D. Quantum fluctuations should be incorporated for the critical physics at the Kondo breakdown QCP, where two kinds of bosonic collective modes will scatter two kinds of fermions, that is, conduction electrons and spinons. Gauge fluctuations corresponding to phase fluctuations of the hopping parameter χ ij = χe ia ij are introduced to express collective spin fluctuations [15]. Hybridization fluctuations are critical, playing an important role for the Kondo breakdown QCP. Such four field variables lead us to the following effective field theory in the continuum approximation, L ALM = c * σ (∂ τ -µ c )c σ + 1 2m c |∂ i c σ | 2 + f * σ (∂ τ -µ f -ia τ )f σ + 1 2m f |(∂ i -ia i )f σ | 2 + b * (∂ τ -µ b -ia τ )b + 1 2m b |(∂ i -ia i )b| 2 + u b 2 |b| 4 + V (b * c * σ f σ + H.c.) + 1 4g 2 f µν f µν + SN(µ b + ia τ ), (3) where g is an effective coupling constant between matter and gauge fields, and several quantities, such as fermion band masses and chemical potentials, are redefined as follows λ → -µ b , (2m c ) -1 = t, (2m f ) -1 = Jχ, µ c = µ + 2dt, -µ f = ǫ f + λ -2Jdχ. Fermion bare bands ǫ c k and ǫ f k for conduction electrons and spinons, respectively, are treated in the continuum approximation as ǫ c k ≈ -2dt + t(k 2 x + k 2 y + k 2 z ) and ǫ f k ≈ -2Jdχ + Jχ(k 2 x + k 2 y + k 2 z ). The band dispersion for hybridization can arise from high energy fluctuations of conduction electrons and spinons. | cond-mat.str-el |
Some exotic snowflake silicon micro- 16 and nano-17 structures were observed in the experiment. The proposed nanoflowers could serve precursors in formation of the structures. 16,17 For example, the 5-petaled "flower" in the center of Fig. 1 17 has a very similar structure to Fig. 2c. A combination of metallic [9][10][11] PPs with semiconducting ones, with different band gap widths and types of conductivity, around a central IQD in one SiNF can serve as a background in the developing of a large variety of nanoelectronic devices. | cond-mat.mtrl-sci |
The Fortran 77 libraries Lapack and Arpack are required, but as these are available on a wide range of platforms, portability should not be affected. OpenFCI is released under the Gnu General Public License (Gnu GPL) [5] and is documented using Doxygen [6]. As an open source project, the code can freely be used and modified. The article is organized as follows: In Section II, the FCI method is introduced in the context of the parabolic quantum dot, where we also discuss the reduction of the Hamiltonian matrix by means of commuting operators and configurational state functions. In Section III we discuss the effective two-body interaction. As the technique is likely to be unfamiliar to most readers outside the nuclear physics community, this is done in some detail. | cond-mat.mes-hall |
The dependence of G on J for the parallel configuration is qualitatively similar to that in the antiparallel configuration, therefore it is not shown here. In Fig. 4 the dependence of the critical exchange coupling J P c on the spin polarization of the leads p in the case of symmetric Anderson model and parallel magnetic configuration is shown for three different values of the tunnel coupling Γ. First of all, the critical coupling J P c decreases with decreasing the coupling strength Γ. Moreover, J P c also decreases with increasing the spin polarization of the leads. For p → 1, J P c tends to zero, as only spins of one orientation are coupled to the leads and the Kondo effect becomes suppressed. | cond-mat.mes-hall |
In contrast, below line D it becomes energetically favourable to unload the QD if it is occupied and sufficiently coupled. Note that until now, we have considered only the DC part of the applied voltage. In Fig. 2 the rf-modulation of the left barrier is taken into account by a parallel vertical shift of the three lines in both directions. The magnitude of the shift is V of the QD from the source. Due to the additional AC component loading the QD is possible even for V DC L and V R voltages down to D -during one part of the cycle. | cond-mat.mes-hall |
4 We employ a model where the polarizers and two arms of the ring are described by a single scatterer each, completely analogously to the normal metal ring case in Ref. [44,45]. The modes are labeled by transversal momentum and valley, while the electrons can propagate in both directions, inside both the left and right lead. 46 The amplitudes of outgoing modes, O ≡ OL OR , and incoming modes, I ≡ IL IR , are connected by a scattering matrix S, with O = SI. The important submatrices of S are t and t ′ , given by O R = tI L and O L = t ′ I R , where I L /I R are columns of amplitudes of incoming (into a scatterer) modes from the left/right, and O L /O R are columns of amplitudes of outgoing modes from the left/right, see Fig. 6(b); thus t and t ′ are M × M matrices, where M is the number of modes in one lead. | cond-mat.mes-hall |
1 (b) shows the photo-energy dependence of the surface contribution obtained by the above procedure. The surface contribution is reduced significantly as the photon energy increases. Therefore, we can deduce that the previous VUV-PES spectra of US are strongly affected by the surface states. From the photonenergy dependence of the S 2p core-level spectra, it can be expected that the surface contribution of the valenceband spectrum of US at hν = 800 eV is suppressed by less than 10 %, because the photoelectron energy for the valence bands is 160 eV higher than that for the S 2p core levels. Next, we performed valence-band SX-PES experiment in order to extract bulk dominant electronic structures. Figure 2 shows the photon-energy dependence of the angle-integrated SX-PES spectra of US. | cond-mat.str-el |
In these experiments a 1/10 of the actual voltage has been applied to the input channel. The output characteristics of the Ti/sapphire laser were: 800 nm, 1 ps duration and 10 µJ per flash power, 1 kHz or 200 Hz repetition rate. Note that NTCI(A) does not absorb at 800 nm in contrast to 532 nm utilized in the previous work. The signals were recorded by an averaging Tectronics oscilloscope (512 traces) recording up to 2 GS/second. Figure 1A illustrates the putative electron transfer reactions measured by the capacitive probe upon illumination of our samples. The energy levels of silicon and NTCDI molecules are aligned accordingly to the previously established fact that the light is absorbed by silicon (Figure 1A, the left panel). | cond-mat.mtrl-sci |
1(a)), the pair of incommensurate peaks observed around (0.5±0.02 0 1.73) are well-defined, which indicates the LT-ICM phase. The imbalance in the intensity of the two peaks results from the domain distribution or magnetic structure factor. As the applied pres-sure is increased to p = 0.99 GPa, the incommensurate peak becomes weak and a commensurate peak appears at Q = (1/2 0 7/4) (denoted by white dashed lines in the figure ), where the LT-ICM and CM phases coexist. At p = 1.25 GPa (see Fig. 1(c)), the commensurate peak is dominant, with a minor incommensurate peak, suggesting that the volume fraction of the CM phase increases with increasing pressure. Figures 3 (a) and (b) show the integrated intensities of the magnetic peak for the HT-ICM, CM, and LT-ICM phases as a function of temperature, taken at three different hydrostatic pressures. | cond-mat.str-el |
If T S < T SPT , we would have two separate second-order transitions, T SDW = T S < T SPT , as in the case of the 1111-family. For the 122-family, which has a shorter Fe-Fe bond length, it is expected this would enhance the spin exchange J, likely leading to T S > T SPT . But the SDW will not form before the SPT, since there is no spin exchange until the SPT obtains. So there is only one first-order transition, T SDW = T SPT . Furthermore, this anisotropic Heisenberg model has also been proposed on experimental grounds 16 to fit the spin-wave spectrum seen in the inelastic neutron scattering data. Our theory gives a direct explanation for the observed anisotropy of magnetic exchanges. | cond-mat.str-el |
Therefore, only phonon modes with m 2 = 1 give non-zero contribution to spin-flip transitions (this is an additional reason why we do not need to consider higher phonon modes with m 2 > 1). Thus, only BM-phonons are responsible for the spin relaxation, whereas TM-and SM-phonons (with m 2 = 0) cannot flip the spin. In the framework of Bloch -Redfield theory, 37 the spin relaxation time induced by BM-phonons is given by 1 T 1 = 2π L ∞ -∞ dq(2N ω + 1)|M ω | 2 δ ω 0 - c S R √ 2q 2 = 2π L ∞ -∞ dq(2N ω + 1)|M ω | 2 1 2 3/4 √ c S Rω 0 [δ(q -q 0 ) + δ(q + q 0 )] = 2 5/4 πL 2 √ c S Rω 0 (2N ω0 + 1) |M ω0 | 2 , (53)where ω 0 = |E κ + 0 ,k0,+1/2 -E κ - 0 ,k0,-1/2 |/ ≈ |ω Z -2τ 3 ∆ curv / |, q 0 = √ 2ω 0 /c S R, and N ω = [exp( ω/k B T ) -1] -1 is the Bose distribution function. Note that pure dephasing 1/T ϕ = 0 for BM phonons and 1/T ϕ = O ∆ 4 SO for SM and TM phonons, therefore, 1/T 2 = 1/2T 1 + 1/T ϕ = 1/2T 1 in first-order perturbation theory. We used the Markov and the secular approximations in the derivation of Eq. ( 53). | cond-mat.mes-hall |
Further measurements of the noise spectrum and its temperature dependence, similarly to those carried in superconducting devices 42 will be instrumental in pointing at the correct mechanism. a biased dot configuration. Taking the X-axis (Z-axis) along the electric (magnetic) field, the two-electron orbital Hamiltonian is given by H orb = i=1,2 H SP i + C(r 1 , r 2 ) (B-1) where C(r 1 , r 2 ) = e 2 /κ|r 1r 2 | is the Coulomb interaction between the two electrons, and the single-particle Hamiltonian is H SP i = 1 2m p i - e c A(r i ) 2 + ex i E + V (r i ). (B-2) The double-dot confinement potential is modeled using a quartic potential in the xy plane and a finite potential in the z direction V (r i ) = V xy (x, y)V z (z) V xy (x, y) = mω 2 0 2 1 4a 2 (x 2 -a 2 ) 2 + y 2 (B-3) V z (z) = 0 |z| ≤ L z /2 V z |z| > L z /2 where we consider a much stronger confinement in the z direction, appropriate for typical gate-defines QD structures. This enables us to perform separation of variables in the lateral and z directions, and to approximate the Coulomb interactions using 2-D integrals. The matrix elements of the orbital Hamiltonian are found by adding and subtracting the harmonic potentials centered at ±a, 27 thus we have H orb = h -a (r 1 ) + h a (r 2 ) + W -(r 1 ) + W + (r 2 ) + C h ±a (r) = 1 2m p - e c A(r) 2 + mω 2 0 2 (x ∓ a) 2 + y 2 + eEx (B-4) W ± (r) = mω 2 0 2 x 4 4a 2 - 3x 2 2 - 3a 2 4 ±2xa Using the orthonormalized single-particle orbitals, ψ ±a = N (φ ±agφ ∓a ) we find the single-particle energies and tunnelings in ω 0 units: ǫ ± = ψ ±a |h ±a + W ± |ψ ±a = ǫ 0 + ǫ E 1 ± ǫ E 2 ǫ± = ψ ∓a |h ±a + W ± |ψ ∓a = ǫ 0 + ǫ E 1 ∓ ǫ E 2 (B-5) t = ψ ±a |h 0 ±a + W ± |ψ ∓a = t 0 + t E with ǫ 0 = b + 3 32b 2 d 2 + 3 8 s 2 1 -s 2 1 b + d 2 ǫ E 1 = ∆x 2 5s 2 -2 4(1 -s 2 ) + 3 8bd 2 + ∆x 2 8d 2 ǫ E 2 = ∆xd(1 -g 2 ) 1 -2sg + g 2 1 - 3 4bd 2 - ∆x 2 2d 2 (B-6) t 0 = - 3s 8(1 -s 2 ) 1 b + d 2 t E = - 3s ∆x 2 4(1 -s 2 ) . | cond-mat.mes-hall |
In differential conductance vs. bias measurements we observe a pronounced zero-bias peak for a QPC conductance ∼ 0.8 × 2e 2 /h, which weakens as the conductance increases to 2e 2 /h, and completely disappears above the first plateau. This behavior might indicate that the structure below the first plateau is related to Kondo-like effect [6]. Besides, at T= 70 mK another plateau-like structure at ∼ 1.7 × 2e 2 /h appears. All features observed in this Hole transport in p-type GaAs quantum dots is also explored. Two quantum dots were fabricated with AFM lithography -one rectangular (Fig. 2(b)) with lithographic dimensions 430 × 170 nm 2 , and the other circular (Fig. | cond-mat.mes-hall |
It is easy to check that the introduced quantities S and S À are also Hermitian [under the condition of time homogeneity It 1 It 2 f t 1 À t 2 , as was assumed in the derivation of Eqn ( 4)]. Formula (5) leads to the well known expression for the spectral density of fluctuations in an equilibrium conductor [10]: SO 2G" hO 1 2 1 exp" hOak B T À 1 ! X6 This means that at zero temperature the fluctuations should be proportional to frequency, which is usually interpreted as an analog of zero (vacuum) oscillations in an electromagnetic field. However, as is known from optical measurement, normal photodetectors do not record zero oscillations, because the energy required to excite an atom in the detector cannot be extracted from the vacuum (see, e.g., Ref. [11]). At the same time, zero-point oscillations can be observed (although by a more complicated way than for usual fluctuations) in the Lamb shift of levels [12], in the Casimir effect [13], or with the use of the so-called Mandel quantum counter [14], which is initially prepared in an excited state and hence can record zero oscillations. | cond-mat.mes-hall |
The bottom, cross-shaped metallization (30 nm Ti/40 nm Pt/200 nm Au) was patterned using standard negative lithography, metal evaporation, and metal lift-off. The BCB dielectric (Cyclotene 3022-46, Dow) was deposited using two consecutive spin-coat depositions and soft cures in a vacuum oven, resulting in a final thickness of approximately six microns. The top ELC metallization layer structure and patterning used the same process as that of the cross. The two unit cells (one for each layer) are shown as the insets to the bottom panel of Fig. 4. The MM absorber sample was examined experimentally using a Fourier Transform Infrared (FTIR) spectrometer. | cond-mat.mtrl-sci |
(2 Ṽ = - dθ dt -Im[ ∂φ n ∂R /φ n ] - 2 m 0 Im[ ∇φ n φ n ] • ∇θ. (2.20) Equation (2.18) indicates that θ is a function which is not dependent on t explicitly, leading to dθ dt = ∂R ∂t ∂θ ∂R = ε ∂θ ∂R . (2.21) Therefore, in our approximation to suppress the terms of O(ε 2 ), equation (2.20) is reduced to Ṽ = -Im[ ∂φ n ∂R /φ n ] - 2 m 0 Im[ ∇φ n φ n ] • ∇θ. (2.22) Thus we have obtained equations which θ and Ṽ should satisfy so that the conditions of regularized standard state and Hamiltonian are fulfilled. (c) Driving potential for fast-forward Here we obtain the driving potential V F F to fast-forward the regularized standard state Ψ (reg) 0 . Such limit are taken as ε → 0, α → ∞ and αε ∼ 1 in the fast-forward. | cond-mat.mes-hall |
So for mesoscopic systems if we can bypass the determination of the internal details and the exact Hamiltonian of the system, by using the S-matrix then the S-matrix includes the effect of electron-electron interaction exactly. The approach proposed in this paper is due to some recent experiments [7,8] that are motivated by the possibility of obtaining important information from the scattering phase shift. So the present work is an effort to identify information that can be obtained from such experiments. A series of experiments [9,10] also tell us that resonances in such a system as that schematically shown in Fig. 1 are Fano resonances. Recently it has been shown for some particular potential (namely a delta function potential) in a quantum wire, that at the Fano resonance the density of states (DOS) and partial density of states (PDOS) can be determined exactly from semi-classical formulas involving scattering phase shift, although Fano resonance is a purely quantum phenomenon [11]. | cond-mat.mes-hall |
Graphene is a promising material [1,2] to investigate mesoscopic phenomena in two-dimensions (2d). Unique electronic properties, such as massless carriers, electronhole symmetry near the charge neutrality point, and weak spin-orbit coupling [3] makes graphene interesting for high mobility electronics [4,5], for tracing quantum electrodynamics in 2d solids, and for the realization of spin-qubits [6]. Whereas diffusive transport in graphene and the anomalous quantum Hall effect have been investigated intensively [7,8], graphene quantum dots are still in their infancy from an experimental point of view [9]. This is mainly due to difficulties in creating tunable quantum dots in graphene because of the absence of an energy gap. Also phenomena related to Klein tunneling make it hard to confine carriers laterally using electrostatic potentials [10,11]. Here we report on Coulomb blockade and Coulomb diamond measurements on an etched graphene quantum dot tunable by graphene side gates [12]. | cond-mat.mes-hall |
Our case is quite different than those mentioned above. SnO 2 has a rutile structure in which each Sn is surrounded by an oxygen distorted octahedron. Also, SnO 2 shows very large magnetic moments (∼ 20 µ B , for Mn-doped SnO 2 ) when doped with TM. However, we will show that it is possible to induce magnetism in SnO 2 without doping of TM. Most importantly, the previous studied materials did not consider the magnetic coupling between the defects. We study in detail that coupling to prove that SnO 2 has magnetism without TM doping. | cond-mat.mtrl-sci |
Hence, inter-valley scattering is rarely considered in investigating the transport properties of graphene systems 24 because of the large separation of Dirac points in momentum space. Moreover, valley contrasting physics is also absent in low-energy quasiparticle's tunneling features in this system. However, our result shows that, although low-energy tunneling in monolayer (bilayer) graphene is sufficiently described by massless (massive) fermion approximation, in high-energy limit there will be a significant deviation between this approximation and the real system 18 . With the increasing of barrier height, inter-valley scattering can be realized and amplified significantly as shown in Figs. 6 and8. At the same time, we find that valley discrepancy in the tunneling problem would be gradually magnified by strengthening the external potential. | cond-mat.mes-hall |
Consider any pair of momenta, k and k + π; these two pseudospins feel opposite z fields and the same x field, so they evolve as 2)) and exact dynamics (red, Eq. ( 9)) simulated using light-cone methods [5]. S x k+π (t) = S x k (t), S y,z k+π (t) = -S y,z k (t), The revival in a) at t 20 can be modeled by Eq. (10). where individual spin polarizations for -π/2 < k < π/2 evolve in the same way as in the original problem (7). After the states are doubled, the net y polarization k S y k taken over the extended range -π < k < π vanishes at all times, while the net x polarization k S x k does not change, and so we can then add the interaction S y k S y k ′ to the original Hamiltonian (7) without changing the dynamics. | cond-mat.str-el |