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With X-FELs, ankylography will enable 3D structure determination of a biological specimen from a single pulse before the specimen is destroyed 25,30 . In materials science, ankylography can in principle be applied to investigating timeresolved 3D structure of disordered materials using X-FELs. METHODS SUMMARY 3D image reconstruction in ankylography proceeds by means of an iterative algorithm. A random phase set defines an initial input and iterates back and forth between Fourier and real space where physical constraints were applied. In Fourier space, the magnitudes of the Fourier transform on the Ewald sphere were set to the measured values. The data points not on the Ewald sphere were initially set as unknowns and updated with each iteration of the reconstruction. | cond-mat.mtrl-sci |
The external field is set to zero and the number of spins in each bath are taken to be equal with N = 100. The system bath couplings given in the inset are dimensionless (g i / g 2 1 + g 2 2 ). relaxation of the spin expectation values to equilibrium. P(t) = 1 2 + 3 i=1 S i (t) 2 . (18) The result is plotted in Fig. 3. | cond-mat.mes-hall |
APPENDIX C In this Appendix we illustrate the transfer-matrix method in application to the network model on triangular lattice. Consider, by analogy to Ref. in Fig. 8 are connected to the corresponding top links by dashed lines, reflecting the fact that these links must be identified with each other in order to impose a periodic boundary condition 6 . Scattering of waves at each node is described by the matrix S = S △ , Eq. ( 7), which relates the amplitudes (o 1 , o 2 , o 3 ) to (i 1 , i 2 , i 3 ). | cond-mat.mes-hall |
2(b),(c), we show a snapshot of the resulting tively. The colors indicate the local bond-angle order parameter which is defined to vary from 0 (blue) in the square lattice to 1 (red) in the rhombic [6]. It is clear from the particle position snapshots, Fig. 2(b),(c), that the product nucleus is isotropic for large temperatures and highly anisotropic for small temperatures. We identify the isotropic nucleus with a ferrite and the anisotropic one with martensite [6]. This identification is reinforced by showing that the latter is twinned. | cond-mat.mtrl-sci |
4 the fitting parameters for the low temperature region are considerably changed, see Table 4. The reduction of the Curie constant C and Weiss temperature (θ = 0.22 K), as well as an increase of the temperature independent contribution can be related to the approach of the system to the LFL regime with a more ballistic motion of the 4f electrons and indicate the Kondo effect in the magnetic susceptibility data. In this respect a Curie-Weiss description well within the Kondo regime, i.e. at T T K , despite successful, may appear not appropriate. However, strong ferromagnetic correlations, as indicated, for instance, by a large Sommerfeld-Wilson ratio for YbRh 2 Si 2 [18] and YbIr 2 Si 2 [9], dominate the magnetic susceptibility and may lead to this Curie-Weiss behavior. The reduction of the Curie constant can also be observed experimentally when comparing the magnetic susceptibility per Yb ion of YbRh 2 Si 2 with Y 1-x Yb x Pd 3 (x = 0.6%) where the 4f electrons are not hybridized with the conduction electrons [19]. | cond-mat.str-el |
1 does not allow choosing one of the theoretical predictions as the correct one. The ensemble of points are closer to the Beeman-Tsu relationship, consequently we think that, if one single relationship exists between ( 2 / Γ ) and θ Δ (i.e. a relationship that is independent of the particular microstructure), it is closer to that of Beeman-Tsu than to that of Vink. This conclusion is reinforced by the fact that four of the six theoretical calculations discussed in the previous subsection deliver values of ( 2 / Γ ) that are very close to eq. ( 3). In view of the coincident slopes of Vink's and Beeman-Tsu's formulas (Fig. | cond-mat.mtrl-sci |
( 15) are shown in Fig. 7 as a function of the strength of the scattering events. Recalling that the number of the elementary rings in a network to realize M QW steps is proportional to M 2 , one might expect the overlaps F k measuring the fault tolerance of the devices to scale also as M 2 . However, choosing a certain level of any of these measures (e.g. F 2 = 0.9), we find that the D values corresponding to this level do not decrease quadratically as the size of the networks increases. Instead, this decrease is even weaker than a linear dependence, thus the stability of the devices scale with size in a promising way. | cond-mat.mes-hall |
This is going to make the many-body physics of Frenkel excitons far simpler than the one of Wannier excitons, the "in" excitons (Q 1 , Q 2 ) and the "out" excitons (Q ′ 1 , Q ′ 2 ) of a scattering process being simply linked by Q ′ 1 +Q ′ 2 = Q 1 +Q 2 , due to momentum conservation, without any additional degree of freedom. By noting that the creation of an exciton Q and the creation of an excitation on site n are linked by B † Q = 1 √ N s Ns n=1 e iQ.R n B † n (2.17) B † n = a † n b † n = 1 √ N s Q e -iQ.R n B † Q (2.18) while due to the lattice periodicity, we do have Q e iQ. (R n ′ -R n ) = N s δ n ′ n (2.19) n e i(Q ′ -Q).R n = N s δ Q ′ Q (2.20) it is easy to show that the neutrality term V intra in Eq. (2.4) reads as -ε 0 Q B † Q B Q while the transfer term V trans reads as Q V Q B † Q B Q . Consequently, the free exciton Hamiltonian can be written as H (0) X = ε e n a † n a n + ε h n b † n b n + S X (2.21) S X = Q ζ Q B † Q B Q (2.22)where we have setζ Q = V Q -ε 0 (2.23) Note that, while exciton operators can be used to rewrite the Coulomb part S X of this free exciton Hamiltonian, this is not possible for the kinetic energy parts which still reads in terms of fermion operators. Also note that, although coming from two-body Coulomb interaction, S X appears as a diagonal operator in the Frenkel exciton subspace. | cond-mat.mes-hall |
This conversion differs however from the inverse spin Hall effect in that it is fully coherent and it couples different spin polarization. Measuring this charge current/voltage allows to extract the spin conductance of the cavity. We consider two types of measurements, defining two different spin conductances G (s1) and G (s2) . In the first one, the voltage on terminal 1 is set such that I (s) 1 is a pure spin current, without charge component, I 1 = 0. Then I 2 = -I 3 are entirely due to the conversion of I (s) 1 into a charge current, and the spin conductance of the cavity is defined as G (s2) j ≡ I j /|µ (s) |, j = 2, 3. In the second scenario, I (s) 1 is a polarized current, accompanied by a net injection of charges into the dot, I 1 = 0. | cond-mat.mes-hall |
The latter is elaborated in Appendix A 2, and the reader may convince herself that both approaches yield the same wave functions. III. THE QUASIELECTRON OPERATOR P(η) With the preliminaries in place, we are ready to address the central issue of this paper -how to construct an operator that directly creates a quasielectron localized at a specific point η. In order to present the basic ideas of our construction in the simplest possible way, we will first discuss the Laughlin states. The generalization to the hierarchical states involves several technicalities and will be deferred to Section V. In order to motivate our construction, we start by discussing the qualitative physical picture, before putting it into a formal language. First, recall the case of quasiholes. | cond-mat.mes-hall |
()3 This transformation shifts the phonon operators by a quantity α, while the electron operator is transformed into a new fermionic one associated to a quasi-particle, called polaron, with energy E p . It can be shown that α 2 is the mean number of phonons in the polaron cloud. By applying the LF transformation H0 = D † H 0 D, the atomic Hamiltonian H 0 = ω 0 a † a -gσ z (a † + a) becomes H0 = ω 0 a † a + E p /2,(4) the eigenvalues E n = ω 0 n + E p /2 correspond to the two-fold degenerate eigenvectors |ψ j n , j = D|n, j = c † j |n were the index n = 0, . . . , ∞ refers to the photon number, j = 1, 2 to the electron site number and c † j is the polaron creation operator c † j = Dc † j D † = c † j exp{(-1) j α(a † -a)}. | cond-mat.str-el |
Air flowing around a falling crystal tends to increase its growth via the well-known ventilation effect [14,15]. The airflow produces an effective increase in supersaturation where the edges stick out farthest and the resulting flow is the fastest (see Figure 5). In the case of a falling plate, this aerodynamic effect mostly increases the growth of the thin edges of the plate (i.e., the prism facets). With this overall picture in mind, we now consider the case shown in Figure 6. Here we start with a hexagonal plate and assume a small growth perturbation somewhere on the AB facet that breaks the six-fold symmetry and makes v AB greater than the other five facet growth rates. This perturbation could come from a crystal dislocation, a step-generating chemical impurity on the surface, a piece of dust, or perhaps some other mechanism. | cond-mat.mtrl-sci |
For such times we denote the spectral function by A(ω) and it is easy to show that A(ω) = Tr C [A(ω)] where A(ω) is defined in Eq. ( 35). This function displays peaks that correspond to removal energies (below the chemical potential) and electron addition energies (above the chemical potential). The spectral functions of our system are displayed in Fig. 7. At weak bias the HOMO-LUMO gap in the HF approximation is fairly the same as the equilibrium gap whereas the 2B and GW gaps collapse causing both the HOMO and the LUMO to move in the bias window. | cond-mat.mes-hall |
The deconfinement only occurs at the K = 0 point. Unlike the perturbation of the classical dipolar interaction, the quantum perturbation by the transverse field introduces the quantum confinement. The disorder phase adiabatically connects to the paramagnetic phase in the large field limit. Some interesting research directions include the finite temperature properties, the computation of experimentally measurable quantities, and search for new quantum spin liquid in other 3-dimensional systems. CHC is grateful for the discussion with Peter Fulde. Special gratitude is heartily given to Naoto Nagaosa for leading him to this field. | cond-mat.str-el |
Substituting Eqs. ( 7), (8) into Eq. ( 5) and performing integration we finally have the components of wave function Ψ( r, t) = (Ψ 1 ( r, t), Ψ 2 ( r, t), Ψ 3 ( r, t), Ψ 4 ( r, t)) T : Ψ 2 ( r, t) = Ψ 4 ( r, t) = 0,(9) Ψ 1 ( r, t) = de -a 2 /2 4 √ π 3e β/2δL δ L + e β/2δH δ H ,(10) Ψ 3 ( r, t) = √ 3de -a 2 /2 4 √ π a 2 d 2 + (x + iy) 2 (ad + y) 2 + x 2 × × e β/2δL δ L (1 - 2δ L β ) - e β/2δH δ H (1 - 2δ H β ) ,(11)where δ L = d 2 + ih(-γ 1 + 2γ 2 )t/m, δ H = d 2 -ih(γ 1 + 2γ 2 )t/m, β = (ad + ix) 2 -y 2 . (12) Two terms in Eq. ( 10), (11) labelled by L and H indices show the contribution of light and heavy holes correspondingly. In particular the item 1/δ L exp(β/2δ La 2 /2) in Eq. | cond-mat.mes-hall |
It can be solved by computing the equations of motion (EOM) for O k,l-1 and O mn,l-2 , and obtain the set of EOM for the strings from l to the unit lengths. Solving the set of EOM bottom up, the right hand side of Eq. ( 9) becomes a large number of the quantum fluctuations of the string excitations of all lengths from l -1 to unity. For example, in Eq. ( 9) there are l O k,l-1 and l(l -1)/2 O mn,l-2 fluctuations, and they have the equations of motion of l -1 string fluctuations of l -2 length, and (l -1)(l -2)/2 string fluctuations of l -3 length, and so on. One can safely assume that their phases are random so that the inhomogeneous part in Eq. | cond-mat.str-el |
The matrix V is chosen such that (νV ) † = (νV ) R , from which it follows that Ũ R = νV ÕR V T ν † . M DIII is therefore isomorphic to the manifold spanned by O = Õ ÕR , Õ ∈ SO(N ). Defining δ Õ = ÕT d Õ, the invariant arclength and mea-sure can be written in terms of δU DIII = δ Õ + δ ÕR as ds 2 DIII = Tr(δU DIII δU T DIII ) = 4 k<l (da 2 kl + db 2 kl ), dµ DIII (U ) ∝ k<l da kl db kl ,(5)where, in spin grading, δU DIII = da db db -da , da = -da T , db = -db T . (6) The parametrization (6) follows from δU DIII = -δU T DIII and δU DIII = δU R DIII .The arclength ds 2 DIII and the measure dµ DIII (U ) are invariant under δU DIII → W δU DIII W T , with W ∈ O(N ). If W also sat- isfies W R = W -1 , such a transformation preserves the symmetries of δU DIII . In the case of the classes C and CI we omit the spin degree of freedom, and we use N to denote the size of the scattering matrices without spin. | cond-mat.mes-hall |
According to a previous Raman experiment, [15] which suggested that the pressure suppresses the CO phase, we can identify for the 150K peak the melting of CO because this peak is flattened by pressure. The second peak (at 80K) can be identified as the stabilization of the I-FM phase previously observed3. However, the film used in the present study is thinner than those studied previously, and therefore its magnetic signal is completely overwhelmed by the substrate response. Consequently, the magnetic nature of these phases is not clear. Nonetheless, the strong magnetic field dependence of their transport properties suggests that both phases possess a magnetic order. The effect of pressure on the CO phase can be interpreted considering the double exchange theory, since the application of external pressure would increase the Mn-O-Mn angle toward 180°. | cond-mat.mtrl-sci |
The work done by the fundamental mode on the atoms during these hops is released irreversibly and leads to dissipation. We find that the activation energies associated with the hopping rates are of the same order of magnitude as those extracted from Fig. 4. Another observation is that the fitted damping rates in Fig. 4 do not seem to depend on the bar size. This invariance is another indication of the quality factor being limited by defects near the geometric edge regions. | cond-mat.mtrl-sci |
In particular, an additional Fermi liquid renormalization factor for the Wilson ratio has the form: (1+G1) (1+F0)(1+G1)-(G2) 2 , where F 0 corresponds to the usual spin density-spin density interaction, and G 1 and G 2 represent additional spin-orbit coupling-induced quasiparticle interactions that vanish in the absence of the spinorbit coupling 23 . Given that the nature of the underlying quasiparticle interactions among spinons is quite different from that between electrons, it is conceivable that the Wilson ratio in metal and the spin liquid can also be quantitatively quite different. Since the interactions between spinons are generally believed to be much stronger, we may expect a bigger Wilson ratio in the spin liquid phase. Quantitative estimation would require the derivation of the full Fermi liquid interaction function, which may be an excellent topic of future study. Acknowledgments We thank R. S. Perry and H. Takagi for showing us unpublished data on Na 4 Ir 3 O 8 and many inspiring dis-cussions. We are also grateful to Arun Paramekanti, Nic Shannon, Leon Balents, and Mike Norman for sharing their insight with us. | cond-mat.str-el |
We are grateful to the Spanish agency CICYT for financial support trough the projects MAT2005-3866 and MAT2007-66719-C03-02 and the Consolider contract (NanoSELECT). Conclusions The adsorption of NO on a flame annealed cleaned Pt(111) surface from dipping in an acid nitrite solution was studied by near edge X-ray absorption fine structure spectroscopy (NEXAFS), X-ray photoelectron spectroscopy (XPS), low energy electron diffraction (LEED) and scanning tunneling microscopy (STM) techniques. LEED patterns and STM images show that no long range ordered structures are formed after NO adsorption on a Pt(111) surface. NEXAFS and XPS spectra confirm that the main specie adsorbed in these conditions is nitric oxide. The NO molecule adsorbs molecularly on the surface trough the N atom in a defined orientation being the maximum saturation coverage about 0.2 monolayers. Besides the NO molecule, NO 2 , and N 2 species are present on the surface in small amounts. | cond-mat.mtrl-sci |
det[ψ j (x k )] j,k=1. ..N , ψ j (x) being the single particle orbitals for the given external potential. Note that Φ F vanishes every time two particles meet as required by Pauli's principle, and hence describes well the impenetrability condition g → ∞ for the bosons. In our specific case, the orbitals for a ring of circumference L and open boundary conditions areψ j (x) = (2/L) sin(πjx/L)(12)with j = 1, ..., ∞. As a consequence of the Bose-Fermi mapping, all the bosonic properties which do not depend on the sign of the many-body wavefunction coincide with the corresponding ones of the mapped Fermi gas. This is the case e.g. | cond-mat.mes-hall |
Experimentally, internal magnetic fields, anisotropic magnetoresistance (AMR), and isotropic magnetoresistance (IMR) can either destroy quantum interference or mask its occurrence. Difficulties in interpretation can be especially severe in carrier mediated ferromagnets because of the sensitivity of quasiparticle properties to the magnetic microstructure. Theoretically, the role of exchange splitting when combined with intrinsic and extrinsic spin-orbit interactions alters quantum interference in a way which has previously been incompletely articulated. At any rate, it is agreed that neither WL nor WAL survive in clean strong ferromagnets for which the magnetic length l H = ( /(eH int )) 1/2 is smaller than the quasiparticle mean free path l. (Here H int is the internal field of the ferromagnet.) On the other hand, traces of quantum interference are expected to survive when l H is larger than l. Larger values of l H /l can be due either to weaker internal fields or to stronger disorder. Because their moments are dilute and randomly distributed, diluted magnetic semiconductors like (Ga,Mn)As have short mean-free-paths (l 5 nm) and weak internal fields (l H 100 nm in the absence of external fields at 5% Mn). | cond-mat.mtrl-sci |
2: This indicates that our CDMFT results show good convergence to the thermodynamic limit. Such rapid convergence may be attributed to suppressed inter-triangular correlations under the corner-sharing topology of the kagome lattice, as discussed in the localized spin models [26,27]. These characteristic T dependences indicate a hierarchy of energy scale for relevant degrees of freedom. Charge fluctuations are firstly suppressed by large U , being signaled by the suppression of ρ D which includes a doublon-holon pair. At a lower energy scale where ρ S decreases steeply, the spin degree of freedom is frozen out, and finally, the chirality is quenched at the lowest energy scale, corresponding to the suppression of ρ K . We identify the characteristic energy scales for charge, spin, and chirality degrees of freedom by the temperatures T charge , T spin , and T chiral where ρ D , ρ S , and ρ K are suppressed most rapidly (the inflection points indicated by the triangles in Fig. | cond-mat.str-el |
The discovery of superconductivity in LaFeAsO 1-x F x with T c = 26 K [1] has attracted interest due to structural and magnetic similarities with high-T c cuprates. To date, much effort has been devoted to the search for new iron-based compounds exhibiting an even higher T c . By replacing La with other rare earths, such as Sm [2,3,4], Ce [5], and Nd [6], T c has been raised to 55 K for Sm and to 54 K in the oxygen deficient RFeAsO 1-δ systems (R= Nd [7], Gd [8]). In both cases, the magnetic and structural transitions in the undoped parent material are suppressed before entering the superconducting phase. Further studies have shown that the ternary FeAs compounds AFe 2 As 2 (A=Ba, Sr, Eu, and Ca) share similar magnetic and structural properties as the RFeAsO parent compound [9,10,11,12,13], and exhibit superconductivity by doping A with K or Na [10,14,15,16,17] or by applying pressure [18,19,20] to suppress the magnetic and the structural anomalies. These similarities suggest that the physics of both families of materials is dominated by FeAs layers and that 'intercalated' layers serve primarily as tunable charge reservoirs. | cond-mat.str-el |
This is because the conductance is changed noticeably near T = 0.7, and then the shot noise through the J (2) channel changes abruptly as well. We did not address the amplitude of ∆G. As pointed out by Kang [10], the asymmetrical structure of the QPC induces an larger dephasing rate in the experiment. In this case, the dephasing rate depends on not only ∆T but also on the change of the phase shift through the QPC, which requires additional information from experiments, such as measurements in the device setup in Ref. [4]. In conclusion, we have discussed the dephasing mechanism due to charge fluctuations of a quasi-bound state in a quantum point contact. | cond-mat.mes-hall |
For d ≥ 2, however, the gap ǫ = O(L -d ) vanishes faster than the wavevector spacing π/L. Thus, as L → ∞, the divergence in the bond number equation is confined to the zero mode and the stucture of the equation mimics that of so-called "sublattice-symmetric spinwave theory" [14]: 1 N q λ ω q → 1 N ǫ + c = 1 + 2S. (14) The details of the lattice enter as a single geometrical constant, c . = 1.393 (square), c . = 1.1567 (cubic), etc. The staggered moment extracted from the large separation limit of X ij = χ 0 ij is M = S -1 2 (c -1) , the usual spinwave result [15]. | cond-mat.str-el |
I. INTRODUCTION Applications of graphene with its unique physical properties 1,2,3,4,5 in nanoelectronics 6,7 , magnetism and spintronics 8,9,10,11,12 , hang crucially on its bandgap and spin ordering at the zigzag edges. A bandgap can be opened in graphene by breaking the certain symmetries. For example, interaction of graphene with its substrate, such as SiC, leads to the charge exchange between them which breaks the sublattice symmetry 13 . Moreover, the quantum confinement effect also has been found to introduce a small bandgap in graphene nanoribbons 14 , just as was predicted earlier theoretically 15,16,17,18 . The effect of bandgap opening and spin ordering between the zigzag edges are found to be directly linked to each other 20 . | cond-mat.mtrl-sci |
To achieve consistency we multiply both sides with G ↑ (iω n )G ↓ (iω m-n ) and G ↑ (iω n ′ )G ↓ (iω m-n ′ ) and sum over fermionic Matsubara frequencies iω n and iω n ′ . We further put m = 0 to keep the effective interaction real. Since the lowfrequency contributions are the most relevant ones, we obtain U χ ee = U χ ee - G ↑ G ↓ L 2 ↑↓ χ ee + G ↑ G ↓ L ↑↓ ,(8)where χ ee = β -1 n G ↑ (iω n )G ↓ (iω -n ) is the static electron-electron bubble. Further on we used a notation L ↑↓ (iω n ) = 1 β n ′ G ↑ (iω n ′ )G ↓ (iω -n ′ )Λ sing ↑↓ (iν -n-n ′ ) ,(9a) G ↑ G ↓ X = 1 β n G ↑ (iω n )G ↓ (iω -n )X(iω n ) . (9b) Equations ( 6) and ( 8)-( 9) form a closed set of relations determining the static effective interaction U and the dynamical vertex Λ sing (iν m ) as functionals of the one-particle propagators G σ and the bare interaction U . They were derived and justified in the critical region of a singularity in the Bethe-Salpeter equation from the electron-hole channel. | cond-mat.str-el |
The fitted quantity σ should theoretically be equal to the applied stress. Here, although close, the fitted value tends to be systematically lower than the applied stress. A posssible explanation is that a fraction of the applied stress is relaxed through free surfaces having X as normal. From the NEB calculations, E * and d * can be determined as a function of stress (Table I). Those are important quantities since they define the energy barrier to overcome to form a stable kink pair. In a non stressed system, the kink pair is never stable and the asymptotic value E * is simply 2F k + W m = 1.94 eV. | cond-mat.mtrl-sci |
2) at the position of the dot. For bend radius r much greater than the interatomic distance, the nanotube band structure is described by that of a locally straight tube, 14 including spin-orbit interaction. [7][8][9][10][11] For a nanotube quantum dot 13 of length L r, the effective Hamiltonian to leading order in L/r is H = - 1 2 (τ 3 ∆ SO σ • ŷ + τ 1 ∆ KK ) +g s µ B σ • B ex + τ 3 g orb µ B B ex • ŷ ,(1) where σ i and τ i are Pauli matrices in spin and valley space, respectively, ∆ S0 is spin-orbit coupling energy, and ∆ KK is a valley mixing term due to substrate, contacts, gates, or any disorder that breaks the crystal symmetry. The first two terms describe the two Kramers doublets, while the last term describes the coupling to magnetic fields of spin and orbital moments. Note that orbital moments are always along the nanotube axis unit vector ŷ . We consider only planar devices with magnetic fields applied in plane of the bent nanotube, but this restriction can be readily generalized to bends and fields out of the plane. | cond-mat.mes-hall |
Besides, another important characteristic, the arrival time to the HEL and the plastic wave velocity are in good correlation with experimental data. The main conclusion obtained from these results is that the non-associated anisotropic plasticity model, as it stands, is suitable for simulating elastoplastic shock wave propagation in anisotropic solids. Different HELs are obtained when the material is impacted in different directions; their excellent agreement with the experiment demonstrates adequateness of the proposed anisotropic plasticity model. However, further work is required both in the experimental and constitutive modelling areas to find a better description of anisotropic material behavior and, particulary, Hugoniot stress level. CONCLUSIONS The EOS proposed in this paper represents a physical generalization of the classical Mie-Grüneisen equation of state for isotropic materials. Based on the αβ generalized decomposition of stress tensor, the modified Mie-Grüneisen equation of state combined with nonassociated plasticity model forms a system of constitutive equations suitable for shock wave propagation in single crystals and polycrystalline alloys. | cond-mat.mtrl-sci |
3(b), i.e., the left QD's are singly occupied by the electrons with either parallel or antiparallel spins, then the controlled quantum-state evolution performed by changing the electrostatic potential profile leads to the uniquely determined final quantum states of electrons. This occurs in all parameter regimes I-V (Fig. 5). In the final state, the electrons are localized in either one or two right QD's in the well-defined spin states (cf. the right inset of Fig. 5). | cond-mat.mes-hall |
Usually, the effective Hamiltonian is reached for = ∞. Due to the continuity of the transformation it is readjusted to the flowing Hamiltonian for every value of . The transformation stops automatically when the commutator [H( ), η( )] vanishes which is generically the case for → ∞, i.e., for convergence for → ∞. The structure of the effective Hamiltonian is determined by the choice of the generator η( ). We first choose the generator which leads to an effective model conserving the number of DOs. To this end, we introduce the operator D := i [n i,↑ ni,↓ + (1 -ni,↑ ) (1 -ni,↓ )](4) counting the number of DOs. | cond-mat.str-el |
For example, if M = 3 1 1 3 and K = 2 0 0 2 , we haveM * = 1 1 1 1 , so that, based on the exponent matrices, the electronic and composite-fermion filling factors are given by (ν 1 , ν 2 ) = (1/4, 1/4) and ν * 1 + ν * 2 = 1, respectively. However, based on Eq. (2.5), the composite-fermion filling factors are fixed at (ν * 1 , ν * 2 ) = (1/2, 1/2). Therefore, the matrix M * does not appropriately describe the possible compositefermion filling factors of the system. We would expect that this leads to problematic results, if we used the Chern-Simons approach to obtain a separation between high-energy and lowenergy degrees of freedom. For this reason, we will only analyse the case that M and M * share their ranks. | cond-mat.mes-hall |
Finally, we note that thermalization might also be prevented in nonintegrable systems due to many-body effects, e.g., the presence of a Mott gap. Nonthermal steady states in nonintegrable systems were observed and argued for in recent numerical studies for finite one-dimensional soft-core bosons [10] and spinless fermions [11]. By contrast, thermalization was observed for hard-core bosons on a two-dimensional lattice [14]. Fast relaxation to a nonthermal quasisteady state, socalled prethermalization [29], was recently observed for the fermionic Hubbard model in high dimensions [17]. Further studies of relaxation in nonintegrable many-body systems are therefore desirable, but will not be the subject of this paper. Our goals in the present paper are thus two-fold. | cond-mat.str-el |
Some of the numerical techniques developed here were briefly described in Ref. 7. The spectral function we shall focus on, in the isotropic case, is S(k, ω) = ∞ j=-∞ e -ikj ∞ -∞ dt e iωt 0|S a j (t)S a 0 (0)|0 (2) where S(k, ω) ≡ S aa (k, ω), with a = x, y, or z, and the subscripts on S indicate sites. The tDMRG method calculates the space-time dependent expectation values appearing in Eq. (2) directly, and then one performs the Fourier transforms (FTs) in Eq. (2) to obtain S(k, ω). | cond-mat.str-el |
Note that calculations for zigzag edges have been done for the conductivity using tight binding theory [27], zigzag edges mixing different values of ky,n which strongly complicates the analytical solution. 3 Experimental set-up, samples and shot noise measurement technique Measuring shot noise requires carefully dedicated electronics. There are several ways to detect it such as cross correlation [5,6,7,8,10,11] or SQUID-based resistance bridge [28] techniques. Depending on the nature of the studied system, one must avoid any low frequency noise known as 1/f noise (also called Flicker noise). By measuring the noise spectrum density at low frequency (10 Hz) S I = A I 2 f β where A is the the noise amplitude coefficient and β ∼ 1, we have extracted A ∼ 10 -8 and checked that our set-up is well above the crossover frequency between 1/f and shot noise. In this work, we used a sensitive lock-in detection technique (see also [29,30,31]), to improve the measurement sensitivity. | cond-mat.mes-hall |
We find closed-form solutions for small Péclet number, P 1, in two distinct ranges of θ. For θ c = O(P 1/3 ) < θ < π/4, we show that φ = φ (0) is given by (3.11), and k = k (0) is given by (3.19), or more precisely by (4.1) k (0) ∼ tan θ + C k 0 P, where (4.2) C k 0 = 2C φ 0 tan θ 2C φ 0 Q(m) cos θ + m 2 (1 + l 3 tan θ)f + + m 3 (1 + l 2 tan θ)f - 2C φ 0 Q(n) cos θ + n 2 (1 + l 3 tan θ)f + + n 3 (1 + l 2 tan θ)f - ; Q(p) and C φ 0 are defined by (3.6) and (3.12). Furthermore, (4.3) φ (0) ∼ Čφ 0 P 2/3 , k (0) ∼ Čk 0 P 1/3 0 ≤ θ < θ c = O(P 1/3 ) , where Čk 0 is defined by (3.9), (4.4) Čφ 0 = n 123 4m 123 1/3 f + + f - 2l 123 2/3 , and p 123 (p = l, m, n) is given in (2.26); cf. equations (4.27) and (4.28) in [8]. In effect, we determine mesoscopic kinetic rates, including the attachment-detachment and permeability coefficients in (3.1)-(3.5). We proceed to describing the derivations. | cond-mat.mtrl-sci |
This corresponds to the case that the quantum dot is connected to a normal reservoir via a single mode point contact. A sketch of the system is shown in the inset of Fig. 2. The point contact is assumed to contain a tunnel barrier of transparency Γ. The barrier alone does not mix electrons and holes, therefore its reflection matrix is diagonal in electron-hole space, r = √ 1 -Γ e iξ 0 0 e -iξ = √ 1 -Γ exp(iξΣ 3 ). (32) Here ξ is the phase an electron acquires upon reflection from the barrier. | cond-mat.mes-hall |
The interpretation of the Na-oP8 phase as an "electride" relates it to the Co 2 Si (oP12) structure with Na occupying the cation atomic positions with the interstitially localized electrons occupying empty anion positions. In present work, an alternative approach is suggested for the interpretation of Na-oP8, in which Na atoms occupy the positions of the AuGa structure of the MnP-type (oP8) that is considered as a Hume-Rothery phase stabilized with two valence electrons per atoms. This interpretation has led us to the assumption of Na-oP8 becoming a divalent metal. Summary and conclusions The compressed alkali metals assume complex low-symmetry structures that are quite unusual among elements. The question arises as to what extent these structures are idiosyncratic, how do they relate to the known structural types and whether they follow any known classification (see, for example, Ref. [46]). | cond-mat.mtrl-sci |
For 2D graphene, the semi-classical diffusive conductivity is given by σ = g s g v e 2 h E F τ 2 = g s g v e 2 h k F ℓ 2 , [2] where g s = g v = 2 are the spin and valley degeneracy factors and the mean free path ℓ = v F τ, with the scattering time τ being given at T = 0 by τ(k) = n imp 4π dk ′ V(|k -k ′ |) ǫ(|k -k ′ |) 2 [1 -cos 2 (θ)]δ(E k ′ -E k ), [ 3 ] where V(q) = 2πe -qd e 2 /(κq) is the Fourier Transform of bare Coulomb potential at the transfer momentum q = |k -k ′ | = 2k F sin(θ/2). While the exact RPA dielectric function is known [41], (very dirty). The RPA and Thomas-Fermi results show 2r 2 s C 0 (r s , a = 4d √ πn) as a function of carrier density n. The points of intersection represent the selfconsistent solution, and one can read off n * from the x-axis, n * /n imp from the leftaxis and the "minimum conductivity" σ 0 from the right axis. The inset shows σ 0 as a function of charged impurity concentration showing that (i) it is non-universal, (ii) dirty-samples have σ 0 = 4e 2 /h over a wide range of impurity concentration and (iii) the cleanest samples have σ 0 ≈ 8e 2 /h, but the value is sensitive to the concentration of charged impurities. Also shown is comparison with representative experimental results, where the square shows results from Columbia, diamond from Manchester, and circle from Maryland. These same three samples (with conductivity shown over the full density range) were compared to a high density numerical Boltzmann theory in Ref. | cond-mat.mes-hall |
These low-energy results are sufficient for the discussion of interactions of the electrons with acoustic phonons. Atomic-scale scatterers such as shortwavelength optical phonons can also couple the two K points, 25,26 but for simplicity we disregard intervalley scattering in our discussion of optical phonons below. IV. ACOUSTIC PHONONS IN MONOLAYER AND BILAYER GRAPHENE Long-wavelength acoustic phonons may be treated with continuum theories to a good approximation. An atomistic treatment is required only for optical phonons, see Sec. VI. | cond-mat.mes-hall |
In Section II we explain the analogy with the hydrogen atom. In particular, we consider the conditions when a long range tail of the dipole electric field can be generated by the atom. In Section III we briefly review known properties of an unbound single hole moving in the antiferromagnetic background of the t-J model. Section IV addresses the limiting case of very strong binding. Here we present results of exact diagonalizations for the 4×4 cluster embedded in an antiferromagnetic background. In Sections V and VI we consider the weak binding limit and discuss symmetry properties of the bound states. | cond-mat.str-el |
• The surface roughness initially increases sharply to later on either saturate or increase at a lower rate. • At low temperatures, the pattern wavelength does not depend on target temperature and ion flux, but it increases with ion energy. • Ripple patterns: -Ripples run perpendicular or parallel to the ion beam projection direction depending on the ion beam incidence angle. -Ripples propagate with a non-uniform velocity. -Shadowing effects appear in the ripple temporal evolution for long sputtering times and large θ values. -Ripples are produced by ions with energies ranging from just a few hundreds of eV up to 10 5 eV. | cond-mat.mtrl-sci |
It is clear from Fig. 6c, that analytical calculations are in a reasonable agreement with experimental data and numerical simulations. The presence of damaged surface layer, reported in Ref. [45], and not-measured surface polarization value allow us to consider finite extrapolation length value as a fitting parameter. -20 0 20 Solid curves are analytical calculations based on Eqs. ( 13)-( 16) for fitting parameters L ⊥ =1.3 nm, L z =1.6 nm, different extrapolation lengths λ 1 (0)=0.1 nm and (λ 2 (h) value appeared not important), for SLT; while L ⊥ =0.7 nm, L z =1.4 nm and λ 1 (0)=0.1 nm, λ 2 (h)>>30nm for CLT. | cond-mat.mtrl-sci |
Such behavior is suggestive of a continuous transition. Recent numerical works reported a continuous transition in the weak-coupling regime, 20,33 in disagreement with analytic results 10 as well as previous numerical calculations. 11 However, the analytic results are expected to be reliable in the weakcoupling regime. As was checked by the DBSS procedure, a reliable extrapolation of the gap requires calculations on longer ladders and keeping a significantly larger number of block states than was available in Ref. 20. The same holds for the entropy. | cond-mat.str-el |
I. INTRODUCTION Martensitic transformations and its pairing with ferromagnetism has been a central subject for investigation in the recent years. Especially, some intermetallics show simultaneous occurrence of martensitic and magnetic transitions, suggesting the possibility of controlling the structural transformation by magnetic field and could be exploited for practical applications. Such multifunctional materials are classified under rapidly growing technological field of Ferromagnetic Shape Memory alloys (FSMA). Among the variety of FSMA, Ni-Mn-Ga alloys are a recently synthesized class of alloys that have been studied extensively and hence serve as a reference in the development of new systems [1,2,3]. The stoichiometric Ni 2 MnGa undergoes martensitic transition around 220 K from a L2 1 cubic phase to a low symmetry modulated structure, while the ferromagnetic transition takes place at 370 K [4]. | cond-mat.mtrl-sci |
I. INTRODUCTION Compounds with strong geometrical frustration of magnetic bonds have attracted much attention due to the unusual properties of their manifold ground states and their peculiar spin dynamics. For some systems the degeneracy of the ground state is infinite (in the nearest neighbor exchange approximation), which leads to (i) delayed magnetic ordering and a wide temperature interval with a short-range correlated state, the so called cooperative paramagnet, 1 and to (ii) an enhanced role for the weaker interactions in the eventual formation of a long-range order state at low temperature. Different mechanisms for lifting the macroscopic degeneracy and selecting a specific ground state is a matter of individual consideration for each system. In the case of pyrochlore magnets, (nearly) fully ordered states, cooperative paramagnetic states, spin-glass and spin-ice states have all been observed in different materials. 2 Experimental investigations of excitation spectra often provide the only viable method of determining a specific set of interactions to be taken into account for an individual magnetic compound. | cond-mat.str-el |
3b. This suggests that the nondegenerate ground state has S z = 0, and is an xy plane polarized PFM state, with strong valley mixing. We have further checked a number of system sizes between 24 × 24 to 200 × 200, and found that the xy plane polarized state is always the ground state as long as both L x and L y are commensurate with 3 (that includes all the systems with N e = 12). Otherwise, an Ising PFM state is found to be favorable, as shown in Fig. 3c. This strong systematic finite-size effect can be understood from the graphene band structure, since valley mixing implies order at the wavevector connecting the two Dirac points, and hence period 3 modulations in both lattice directions [19]. | cond-mat.mes-hall |
Isotopic disorder is modelled by replacing with probability 0 < c < 1 the mass of a particle by 1 + δ with δ = ∆m/m. A. Heat transport in isotopically disordered harmonic lattices Heat transport in isotopically disordered harmonic lattices has been thoroughly studied. 10,49,50,51,52 In partic-ular, the case a finite disordered chain connected to two semi infinite perfect chains has been considered, 49 and the corresponding theoretical results were rederived from a different perspective and extended in Refs. 52 and 19. It is suggested in Ref. | cond-mat.mtrl-sci |
The initial amplitude is chosen to be 10% of the bar width. When the bars are "plucked" in the x direction, the x-mode energy is transferred partially over time into the y-direction mode, as evidenced by the center of mass trajectory shown in Fig. 1. The energy transfer between the two transverse modes is attributed to the fact that the silicon dimer atoms on the bar surface {100} align themselves in the diagonal direction as shown in Fig. 2. The dimerized surface of the bar breaks the cubic symmetry of the bulk and, we infer, induces the mixing of the two transverse modes. | cond-mat.mtrl-sci |
Some of our data did suggest polaron peaks, but these polaron peaks do not appear on all films measured, even if grown under identical process conditions. When polaron peaks were evident, the peak-to-peak gap varied little with temperature, and no trend, such as seen in Ref. 4 was apparent. We applied this normalization procedure to dI/dV spectra taken at larger voltage setpoints (V set ≥ 1.0 volt), but found that, in the range of bias voltages where we would expect to see polaron peaks, the signal-to-noise ratio was too low. We also measured atomically smooth films in an applied magnetic field. Shown in Figure 8 are dI/dV spectra measured on a smooth 26 nm thick film grown on NGO in zero field and in a 5 Tesla field. | cond-mat.str-el |
An important step forward was an experiment reported in Ref. 26. Mass-selected clusters of Pt n and Pd n with 1 ≤ n ≤ 15 on Ag(110) were investigated using photoelectron spectroscopy. The authors observed size-dependent d states at binding energies around 2 eV below the Fermi energy. The line widths were observed to be size-dependent as well. From a comparison with total-energy calculations a chainlike shape of the clusters was inferred. | cond-mat.mtrl-sci |
The phase transition of antiferromagnetic ordered state occurs at U c ∼ 1.8 eV. With increasing U , sublattice magnetization m increases, while, the density of states ρ F is almost constant. Due to the effect of the onedimensional potential, the band structure becomes quasi one-dimensional and the Fermi surface nesting is enhanced as shown in Fig. 3. Then, the antiferromagnetic ordered state appears in the presence of the Na order [11]. Figure 2(b) shows m, nn , n yzn zx and ρ F as functions of V with U = 2.5 eV, J = 0.25 eV, T = 20 K and ∆ε d = 0.5 eV. | cond-mat.str-el |
The hidden (Z 2 ×Z 2 ) l symmetry responsible for the hidden order has been found by applying a unitary transformation to the model Hamiltonian. The Haldane gap and 4 l degenerate ground states in an open chain are natural consequences of this hidden symmetry breaking. For an even n = 2l, a periodic chain has a twofolddegenerate dimerized ground state, which preserves SO(2l) symmetry but breaks translational symmetry. These SO(2l) matrix product states with different l are non-Haldane liquid states, which have soliton excitations connecting the two degenerate ground states. However, these SO(2l) matrix product states also contain a hidden antiferromagnetic order characterized by nonlocal string order parameters. Finally, a generalized SO(n) symmetric bilinearbiquadratic model family has been discussed and the ground-state phase diagrams are sketched based on some known exact results. | cond-mat.str-el |
[4]. for our purposes we can use the following simple approximate expression which allows for analytic calculations and provides results that are indistinguishable from the exact results (see Fig 3) proximations. The Random-Phase-Approximation (RPA) is the main approximation used in the present work. The Thomas-Fermi result was derived in Ref. [30] and the "complete screening" result valid for r s a ≫ 1, was obtained by Ref. [5]. | cond-mat.mes-hall |
In previous papers we already show that E1-E2 and E2-E2 channels contribute to the signal. Here we can make the distinction between the different components of each channel. As shown in the bottom of figure 1, we can see, for instance for the (1,1,1)σσ reflection that the main contribution comes from the toroidal moment. This is thus a proof of an anti-ferromagnetic ordering of the toroidal moment in the material. Their axis on each atom can also be given. Nevertheless due to the inversion center, this peculiar current cannot bring macroscopic magnetoelectric comportment. | cond-mat.mtrl-sci |
In the analysis presented below we simplify notations by omitting some indexes, and assuming the summation over repeating indexes. Check of locality of the electron operator ( 8) is obvious, and follows from the commutation rule for phase operators. The statistical phase θ of the operator ψ α is defined as:Using the simple relation and the commutation relation for bosonic phase operators we find that our electron operators (8) are fermions with the phase θ = π. Finally, the total charge at the quantum Hall edge is Therefore, using the relation (A1) we find which means that the fermion (8) in our model has an electron charge, e = 1. The only non-trivial question is whether the condition of the cancellation of the anomaly inflow imposes any constraint on the interaction matrix V ? The answer is no. | cond-mat.mes-hall |
In the right hand panel we therefore show the current measured at time tΓ = 1 (solid lines) and 1.25 (dashed lines) as a function of voltage. At large voltage, one observes a slow increase of I(V ) in the "Coulomb blockade" regime (V U ) followed by a more rapid increase in the current once the voltage bias exceeds the splitting between the Hubbard bands of approximately U . We do not find a rapid increase (comparable to the U = 0 curve) in the current near V = 0, presumably because the time-scales reached in this simulation are not long enough for a Kondo resonance to form, or because the latter is destroyed by even a small applied voltage. However, at voltages V /Γ 2, where the Kondo resonance is wiped out, we expect our hybridization expansion results to be fairly accurate. Figure 17 compares the hybridization expansion results to the mean field and perturbative calculations. The right hand panel shows the comparison to perturbation theory described in Ref. | cond-mat.mes-hall |
Our results were obtained using the Anderson model corresponding to the dynamical mean-field approximation to the single-orbital Hubbard model on a Bethe lattice. 7,19 This model is specified by an interaction strength U , a hopping parameter t, and a carrier concentration n. At carrier concentration n = 1 the dynamical mean-field approximation to this model has a metal-insulator transition at a critical interaction strength U ≈ 5t (at the temperatures we study) 7 and we shall see that the bold expansion behaves very differently in the insulating and metallic phases. IV. RESULTS: METRICS An important metric is the weight p(k) of terms at order k in the expansion for Z. This is shown in Fig. 2 for different interaction strengths U at inverse temperature β = 10/t. | cond-mat.str-el |
This in turn requires calculations involving millions of atoms, well beyond the current capability. This significant challenge is overcome by a Quasi-Continuum reduction of the proposed real-space finite-element formulation, and enables computation of the electronic structure of multi-million atom systems without significant loss of accuracy. This is a multi-scale scheme which facilitates a systematic coarse-graining of electronic structure calculations in a seamless manner that resolves detailed information in regions where it is necessary (such as in the immediate vicinity of the defect) but adaptively samples over details where it is not (such as in regions far away from the defect) without significant loss of accuracy. The real-space formulation of orbital-free density-functional theory, and a finite-element discretization which is amenable to coarse-graining are crucial steps in its development. The approach is similar in spirit to the quasi-continuum (QC) method developed in the context of interatomic potentials (cf. e. g., Tadmor et al. | cond-mat.mtrl-sci |
Structural data were measured by x-ray diffraction on a Siemens D5000 diffractometer upon cooling and warming in the temperature range between 15 K and room temperature using a home-built cryostat. The analysis of this data was carried out by applying the Rietveld technique using the program Fullprof. 43 III. La1-xAxCoO3 WITH A = Sr, Ba A. Thermal Expansion Fig. 1 displays the thermal-expansion coefficients α(T ) of (a) the La 1-x Sr x CoO 3 and (b) the La 1-x Ba x CoO 3 series. | cond-mat.str-el |
( ) is generically represented as ( ). General fusion rules (see Appendix V F) produce new connected terms by connecting two or more connected parts. Such terms are generically represented as ( ) (fusion of two connected parts). For example, all the rules given in Eqs. (24) lie in the generic class of rules,( ) × ( ) = ( ) + ( ) . (26) Note that the notation . | cond-mat.str-el |
The orbital correlation function T ij = - 1 12 at α = 0 may be understood as an average over the orbital triplet (+ 1 4 ) and the two non-dimer bonds (each - 1 4 ). When α increases, this value is weakened by orbital fluctuations, and undergoes a transition at α = 0.32 to a regime where orbital fluctuations dominate, and T ij is close to zero. Above α = 0.69, T ij becomes positive, and approaches + 1 12 as α → 1, indicating that the wave function changes to the staticdimer limit. While T ic vanishes on the c bond here, the cluster average has a finite value due to the contribution T ij = 1 4 from the active non-singlet bond. The spin-orbital correlation function C ij also marks clearly the three different regimes of α. When α < 0.32, C ij has a significant negative value [Fig. | cond-mat.str-el |
However, without these considerations, this result demonstrates that there is no direct relation between the Peierls instability in a system of noninteracting particles and CDWs in real systems. The same point can be made using a linear response approach 20 : an interacting half-filled electronic system is stable against infinitesimal perturbation. Only a finite distortion, and only if e-ph coupling is larger than a critical value, can be stable. This is already a very serious reservation, but nonetheless it is instructive to step back and investigate the "classical" Peierls instability in a noninteracting system. There is no question that the susceptibility of this system is logarithmically divergent, but there are interesting, and important questions left to ask. First, is this divergency robust with respect to small deviations from a "perfect nesting", as is always the case in real materials? | cond-mat.mtrl-sci |
5(b) for comparison. The p -T phase diagram clearly indicates that the HT-ICM and CM phases at ambient pressure become resistant against pressure, and T N1 and T CM slightly increase with increasing pressure. By contrast, the LT-ICM phase at ambient pressure becomes unstable with increasing pressure and finally the system becomes single-phase CM above p ∼ 2 GPa through phase coexistence. Recent study of dielectric measurements under hydrostatic pressure in HoMn 2 O 5 revealed that the weak electric polarization in the X phase grows upon applying pressure [15]. A comparison of our p -T magnetic phase diagram with the p -T dielectric phase diagram [15] reveals that there is a one-to-one correspondence between the pressure-induced commensurate spin state and the pressure-induced electric polarization, which is also seen in the H -T phase diagram [10,11]. This suggests that commensurate magnetism leads to bulk electric polarization. | cond-mat.str-el |
INTRODUCTION Printable electronics presents new promising technology area, which may pave the way to many new low-cost products. Feasible products based on printable electronics might include ultra cheap radio-frequency identification tags, inexpensive and disposable displays/electronic paper, interior interconnections, parts of an electronics assembly (e.g PWB, phone chassis etc..), sensors, memories, and wearable user interfaces. Different materials ranging from metallic inks towards organic materials can be used to produce printed electronic circuits or modules. A key advantage is the ease and speed of fabrication, which simplifies device manufacturing and lowers cost. Ideally, printed electronic circuits can be processed from solution by using similar processes as well-known printing of ink on paper or foils [1]. In addition, as the temperature used in the production process is relatively low, various substrates can be used including cheap, lightweight, and flexible plastic instead of glass [2]. | cond-mat.mtrl-sci |
At cryogenic temperature, the PL spectrum reveals the presence of a series of well-resolved peaks (see Fig. 1a). The main PL emission bands denoted X e1-hh1 and X e1-lh1 are ascribed to the heavy-and light-hole excitonic transitions, respectively. The lower energy transitions, labelled e-Be, results from the recombination of free electrons with Be acceptors [6]. As for the line [Be, X], it originates from the emission of excitons bound to acceptors. The spectra of medium and highly doped samples are more complicated. | cond-mat.mtrl-sci |
The time-averaged current in small Hubbard rings shows a collapse of the currents to a universal curve when the currents are plotted as a function of the Landau-Zener parameter, 32 sharing the same qualitative traits as our Fig. 3, with a negligible current before the breakdown and a linear I-V characteristics at biases larger than the threshold. An important conclusion of our work is the confirmation that the mechanism of the dielectric breakdown corresponds to the Landau-Zener tunneling mechanism and this mechanism survives upon coupling the interacting region to leads. It should be noted that another very recent tDMRG study by Kirino and Ueda 38 has adressed the destruction of the MI state upon application of a strong voltage as well. There are important differences with our work, though. In Ref. | cond-mat.str-el |
The spin value of the cation vacancy depends on the charge state of the defect, with neutral (q=0) and negatively charged (q=-1) indium vacancies forming S=3/2 and S=1 (triplet) states respectively. The effect is quite similar to what has been observed in thin films of liquid oxygen where two out-of surface oxygen orbitals form triplet states, which interact ferromagnetically. 31 These assumptions suggest the following schematic picture. We propose that vacuum annealed In 2 O 3 has both oxygen and indium vacancies. The former act as donors and supply electrons to the conduction band, while the latter act as acceptors with a localized spin. Free electrons from oxygen vacancies will mediate an interaction between the triplet indium vacancies. | cond-mat.mtrl-sci |
With the longitudinal conductivity of 2DEG σ xx = e 2 τ E F /( 2 π), we find the ratio of spin polarization to the current for the 2DEG is S (e) y j x = χ yx σ xx = 2πm e α e eE F . (47) Compared with (46), we find the CISP of 2DEG is inversely proportional to Fermi energy. This means the ratio for 2DEG decreases for heavier doping. This different Fermi-energy dependence stems from the different types of spin orientation for 2DEG and 2DHG. The spin orientation, which is the expectation value of spin operator S for an eigenstate, is given by kµ| Sx |kµ = -S 0 k sin θ + µk 2 S 1 sin θ, (48) kµ| Sy |kµ = S 0 k cos θ -µk 2 S 1 cos θ, (49) kµ| Sz |kµ = 0,(50)for 2DHG, andkµ|S (e) x |kµ = -µ sin θ,(51)kµ|S (e) y |kµ = µ cos θ, (52)kµ|S (e) z |kµ = 0,(53) for 2DEG. In the following, we take S kµ as short for the spin orientation above. | cond-mat.mes-hall |
For a number of technical reasons, efforts to change the diamond process from a molecular-hydrogen based process have been made in the last years. Also, O-and OH-containing chemicals have been also explored in both plasma-assisted and hot-filament techniques. For example, small quantities of oxygen and water vapor have been added to microwave plasma reactors for the purpose of oxygen addition. Small percentages (0.5 to 2%) of oxygen and small percentages of water vapor (0 to 6%) improve the quality of the Raman spectra and decrease the deposition temperature. However, higher percentages of both have been demonstrated to degrade the diamond quality. Therefore, a careful C-O-H relationship is a must if pure diamond phase is to be successfully produced in a CVD-type system. | cond-mat.mtrl-sci |
This quantity is equal to zero when hybridization is absent. Therefore thermodynamic potential is equal toF = F 0 - 1 β U (β) c 0 ,(17) In Fig. 4 are depicted some of the simplest vacuum diagrams. The first three diagrams are of chain type and are originated from the ordinary Wick contributions. The last three diagrams contain the correlation functions and are determined by the new contributions of GWT. The factor 1 n , where n is the perturbation theory order, present in these diagrams makes it difficult to carry out the summation over n. As is usual in such cases [23] we employ a trick, that of integrating over the interacting strength λ. | cond-mat.str-el |
These spin-lattice simulations were done to validate our analytical calculations for the effective anisotropy model. Throughout this work we compared the results of the spin-lattice simulations with H = H ex + H d with the results of micromagnetic simulations. We never found any noticeable difference. We present the results for a disk-shaped and a prism-shaped nanoparticle because these two geometries are the most common ones in experiments. A. Disk-shape nanoparticle Our effective anisotropy approximation provides the exact solution for all homogeneous states for a nanodisk. Therefore we do not need to justify it for the homogeneous states. | cond-mat.str-el |
Therefore we consider Ir 4+ ions on the pyrochlore FIG. 1: Phase diagram based on the slave rotor approximation and strong coupling limit, as a function of Hubbard repulsion U and spin-orbit coupling λ (relative to hopping t). The four main phases occuring for moderately strong electron-electron repulsion are a Metallic phase, Topological Band Insulator (TBI), Topological Mott Insulator phase (TMI), and Gapless Mott Insulator (GMI). The dashed line denotes an additional zero-gap semiconductor state due to an "accidental" gap closing. The dotted line schematically separates the large-U region, where magnetic ordering is expected. As discussed in the main text, long-range Coulomb interactions are expected to induce an excitonic region in the vicinity of the Metal-TBI boundary shown here. | cond-mat.str-el |
As we demonstrate in present paper, this model can be treated numerically using the same MacKinnon-Kramer finite-size scaling algorithm Three links enter a node and three links exit each node; the nodes with S-matrices Eq. ( 7) are depicted as triangles. threshold. This procedure reproduces the exact threshold value and yields a very accurate estimate for the critical exponent. In Refs. 23,24,25,26,27 the classical RG procedure 22 was generalized to the quantum bond percolation. | cond-mat.mes-hall |
To derive the transformational properties under T x and T r2 , we compute the transformation properties of the original spinon matter fields. The procedures for doing so have been described in details in Ref. 7, here we shall just state the results. Let η 1 , . . . | cond-mat.str-el |
Doing the same for the square lattice amounts to extending the result of the ladder calculation to a square lattice. Recalling for the ladder [Eq. (3.8)], the pseudospin Hamiltonian for the two-plaquette unit was H σ = t ′2 V [const(x) + T A • T B ] + t ′2 2V 1 1 + x (T y A + T y B ) (3.9) The similar result for the perpendicular direction in the plane would be H σ = t ′2 V [const(x) + T A • T B ] - t ′2 2V 1 1 + x (T y A + T y B ) (3.10) The minus sign for the single-plaquette terms in the second case is because the bond-ordering in the two perpendicular directions are the pseudospin in +y and -y directions respectively(see Fig. 2). Hence for the infinite square lattice, we get H square = t ′2 V α,β [const(x) + T α • T β ] (3.11) which is the antiferromagnetic Heisenberg Hamiltonian. Since, the square lattice is two-dimensional, its ground state will possess long range order. | cond-mat.str-el |
(17) Here b 1 , b 2 , b 3 and B ef f are fitting parameters. It is pertinent to note that while deriving Eq. ( 17) we use Eq. ( 16) with substitution (B -B ef f ) for T . Then, Eqs. ( 16) and (17) are not valid at B B c0 . | cond-mat.str-el |
As shown by the blue curve of Fig. 6(c) and the lattice parameter profile in Fig. 6(e) a 10% larger value of R i will split the peaks apart while smaller R i values will bring them closer. For the given bilayer tube the fit is sensitive to variations in R i of about 4%. Here one must notice that, as shown in Fig. 5(a), the x-ray diffraction method probes a very reduced volume of the rolled-up crystalline layers. | cond-mat.mtrl-sci |
I/ Model In order to compare experimental data on RHEED rod-spacing oscillations with calculations, we proceed in 3 steps: (1) we show how the finite lateral size of the 2D islands plays on the misfit definition. Moreover, we show that an elastic misfit may be defined even for homoepitaxy since a small 2D island has a different in-plane lattice spacing than in the bulk (section I1). (2) by using the concept of point forces, we formulate the problem of elastic relaxation of coherent 2D epitaxial deposits. The corresponding mean deformation in the islands actually leads to an oscillatory behaviour with coverage (section I2). (3) by using a mean model of a weakly deformed layers pile-up, we calculate the intensity diffracted by such coherently deposited 2D islands and give analytical expressions of the rod-to-rod distance oscillation of the corresponding RHEED pattern (section I3). I1/ Epitaxial misfit The natural misfit m o of two infinite cubic phases A (crystallographic parameter a o ) and B (parameter b o ) is defined for parallel axis epitaxies as:m o =(b o -a o )/a o(1) Thus the parameters a o and b o are linked by b o =a o (1+m o ). | cond-mat.mtrl-sci |
(12) The function N includes the valueε = ε ′ + iε ′′ ,(13) which determines influence of the spatial dispersion on the radiative broadening (ε ′ γ r ) and shift (ε ′′ γ r ) of the doublet levels. ε ′ and ε ′′ are equal 9,20 : ε ′ = Re ε = 2B 2 1 -(-1) m c+m v cos kd ,(14) ε ′′ = Imε = 2B × (1 + δ mc,mv )(m c + m v ) 2 + (m c -m v ) 2 8m c m v -(-1) mc+mv B sin kd - (2 + δ mc,mv )(kd) 2 8π 2 m c m v ,(15) B = 4π 2 m c m v kd [π 2 (m c + m v ) 2 -(kd) 2 ] [(kd) 2 -π 2 (m c -m v ) 2 ] (if kd → 0, ε → 1 (m c = m v , an allowed transition ) and ε → 0 (m c = m v , a forbidden transition)). II. THE TIME-DEPENDENCE OF THE ELECTRIC FIELD OF REFLECTED AND TRANSMITTED LIGHT PULSES With the help of the standard formulas, let us go to the time-representation ∆E ℓ (z, t) ≡ ∆E ℓ (s) = = 1 2π +∞ -∞ dω e -iω s ∆E ℓ (z, ω), s = t + νz/c, (16 ) E r (z, t) ≡ E r (p) = 1 2π +∞ -∞ dω e -iω p E r (z, ω), p = t -νz/c. (17) The vectors ∆E ℓ (s) and E r (p) have the form ∆E ℓ (s) = e ℓ ∆E ℓ (s) + c.c., E r (p) = e ℓ E r (p) + c.c..(18) It is seen from expressions (11) and ( 12) that the denominator in integrands of ( 16) and ( 17) is the same. It may be transformed conveniently to the form ω1 ω2 + i(ε/2) (γ r1 ω2 + γ r2 ω1 ) = (ω -Ω 1 ) (ω -Ω 2 ),(19) where Ω 1 and Ω 2 determine the poles of the integrand in the complex plane ω. | cond-mat.mes-hall |
(41). Higher orders. Diagrams with larger leading orders (typically enclosing several hexagons) as well has further renormalizing corrections to low leading order diagrams are more problematic to obtain using simple arguments. Nevertheless, the calculation can be systematically extended to significantly higher orders using the expansion scheme presented in Section II B 2. Up to the chosen order, this requires (i) the enumeration of all terms appearing in Ĥ, (ii) a careful enumeration of fusion rules that proliferate as the order increases, (iii) the expansion of Ô-1/2 and (iv) the evaluation of Ô-1/2 Ĥ Ô-1/2 . Note that the two last steps explicitly require using the fusion rules obtained in step (ii). | cond-mat.str-el |
Based on a generalized Lieb-Schultz-Mattis theorem, Li 46 studied SO(n) antiferromagnetic models for n = 4, 5, 6. He found that the SO(4) Heisenberg model is gapless, while SO(5) and SO(6) Heisenberg models are suspected to have a gap. Together with our results, we predict that the SO(n) Heisenberg model for n ≥ 5 belongs to the dimerized phase with a finite-energy gap. VI. CONCLUSION In conclusion, we have introduced a class of SO(n) symmetric spin chain Hamiltonians with nearestneighbor interactions, whose exact ground states are two different SO(n) symmetric matrix product states depending on the parity of n. For an odd n = 2l + 1, a periodic chain has a unique ground state, which preserves an SO(2l + 1) rotational and translational symmetries. The SO(2l + 1) symmetric spin chains with different l are directly related to quantum integer-spin chains belonging to the Haldane gap phase with a hidden antiferromagnetic order characterized by nonlocal string order parameters. | cond-mat.str-el |
We first study a three-terminal device by decoupling one of the superconductor terminal (2 or 4). Fig. 1b shows the Andreev reflection coefficients T 11A and T 13A as a function of incident electron energies E. It can be seen that T 11A and T 13A are quite large when the energy E is within the gap (|E| < |∆|) and exhibit peaks at the Dirac points E = ±E 0 and the gap edge E = ±∆. [17] Similar to the usual normal-superconductor junction T 11A and T 13A decay quickly when E is outside of the gap. [18]. Note that when |E| < |E 0 |, the incident electron and reflected hole are in the same band (see Fig. | cond-mat.mes-hall |
A proper polypyrrole coated textile device could represent a possible solution for heating and cooling and for temperature monitoring. As previously told, one of the first commercial textile products incorporating conductive polypyrrole (PPy) was the Contex conductive textile. More recently, Contex-like textiles, with a modified PPy coating have been commercially developed that are more conductive and thermally stable. While imparting electrical conductivity and a dark colour to the substrates, the coating process barely affects the strength, drape, flexibility, and porosity of the starting substrates. For the measurements discussed in this paper, we use a polypyrrole coated PET fabric which was prepared similarly as described previously in [9,10], with raw chemicals purchased from Sigma-Aldrich and used without further purification. Stochiometric molar ratio of organic acid dopant, anthraquinone-2-sulfonic acid to pyrrole-monomer (i.e., 0.33:1) was used to ensure complete doping level. | cond-mat.mtrl-sci |
25% for simple cubic lattices. This discrepancy could be due to non-uniform particle dispersion in the matrix. Our 14 vol% sample was the highest spin castable formulation. As shown in Figure 3a, samples were observed to have lower residual stresses than control SU-8 samples spin cast and patterned with the same lithographic exposure, bake, and development process. This observation is consistent with reports by Renaud on SU-8 composites with silica nanoparticles [11]. SU-8/Diamondoid Nanocomposites Incorporation of diamantane in SU-8 provides the expected reduction in residual stress, as shown in Figure 3b. | cond-mat.mtrl-sci |
To understand these effects, it is instructive to consider each of the transverse modes individually. According to Eq. 6, for a wide graphene sheet, the transverse momentum with respect to the Dirac point is given by k y ≈ nπ/W where n is an integer. We can view each mode as a ray with an angle of incidence on the junction θ = tan -1 [k y (k 2 f -k 2 y ) -0.5 ]. Fig. 4(a) plots the energyresolved local density-of-states (GΣ in G † ) for a pn junction with a transition length of D w =50nm and a built-in potential V pn =0.5eV . | cond-mat.mes-hall |
. ,(15)where i ′ 0 (ε) = π2 -7/2 P + λρ(ε, d/2)e -d/λ , t ′ 0 (ε) = 2hpi ′ 0 (ε),(16) and η α = |h α |/Γ α-. In Eq. ( 15) we kept only those sub-leading terms in the small-p expansion that have a different θ dependence than the leading terms. The case II contributions are obtained by interchanging L and R and replacing ρ(ε, d/2) by ρ(ε -U, d/2). The total torque is found by adding contributions for cases I and II and integrating over ε. | cond-mat.mes-hall |
3. This fieldinduced breaking of graphene spatial symmetry can be observed for non-integer filling of zero LL, when clockand counterclock-wise currents do not compensate each other. To conclude, we have proved that the orbital mechanism is sufficient to explain the zero LL splitting in graphene. The effect occurs in a perturbative non-critical manner and is an intrinsic property of noninteracting fermions on a hexagonal lattice. As a consequence (observable optically), the orbital splitting should not depend on the LL filling factor, unlike the result from other models. At the same time, the many-body and/or disorder effects can amplify the orbital splitting (even for n = 0), induce an additional symmetry breaking, and bring about a nontrivial field and filling factor dependence of the gap observed experimentally [11,12]. | cond-mat.mes-hall |
A single M3 hole was tapped along the axis with a tool steel tap. The 8 metal samples were immersed in acid for 10 minutes and the 2 Torlon samples were immersed for 24 hours. Volume susceptibility measurements were taken before and after immersion following the same procedure used for the larger samples. Fig. 4 shows the change in volume susceptibility before and after acid treatment for each of the samples. The smaller cylindrical samples should have a higher relative surface contamination due to the large surface to volume ratio. | cond-mat.mtrl-sci |
Given this level of agreement, we believe that the atomic model in Fig. 5e provides a reasonably accurate description of the R30 structure. Obviously, the model can be improved using more rigorous image simulations based on electronic-structure calculations. Atomic models of the R18.5 and R14 structures We have not yet imaged the R18.5 and R14 structures with high resolution by STM. Given the rarity of these phases, it is not unexpected that they are difficult to locate. Still, based on the rotations of the graphene sheets relative to the Ir lattice found in the LEED patterns and the knowledge gained from the R0 and R30 structures, we are able to propose the atomic models of the R18.5 and R14 structures shown in Fig. | cond-mat.mtrl-sci |
2. This behavior of the Kondo phenomenon in the presence of ferromagnetic leads is in agreement with that found by other methods, for instance by the equation of motion for the Green functions 37 and also by the real-time diagrammatic technique. 38 The situation changes when the electron in the dot is additionally exchange coupled to the nonmagnetic reservoir. When the coupling is antiferromagnetic and the coupling parameter J increases, the width of the Kondo peak becomes gradually narrower and narrower. The hight of the peak, however, remains unchanged, as can be clearly seen in Fig. 2 for some small values of the exchange coupling constant. | cond-mat.mes-hall |
Since the ratio depends on the antidot potential at the Fermi level and the magnetic field, the noninteracting model cannot provide an explanation of the sample-independent π phase shift. Moreover, in the absence of interactions, both spins should participate in the resonant scattering, contradicting the experimental observation that only the spin species with the larger Zeeman energy contributes to the resonances. 4,8,11 Thus, experimental findings gave strong motivation for a model that takes electron interactions into account. To explain double-frequency Aharonov-Bohm oscillations, models accounting for the formation of compressible rings around the antidot 8 and capacitive interaction between excess charges were introduced 9 . The first model is based on the assumption that there are two compressible regions encircling an antidot, separated by an insulating incompressible ring. Screening in compressible re-gions, and Coulomb blockade, then force the resonances through the outer compressible region to occur twice per h/e cycle. | cond-mat.mes-hall |
It coincides with easy axes experimentally observed for all compounds [6,7,8]. In order to analyze different contributions of the 5f states to the resulting densities of states we have chosen a new local coordinate system with z-axis directed along total moment J. The partial density of states for 4 orbitals with maximal occupation numbers in both jj-and LS-representations are shown in the Fig. 3. Based on the obtained densities of states we can conclude that there are three occupied orbitals and one partially occupied orbital which forms the peak on the Fermi level. The similarity of PDOS shape in both jjand LS-schemes confirms the applicability of intermediate coupling scheme. | cond-mat.str-el |
Once the Coulomb interaction is included in the Hamiltonian, no analytic solution is available. Several methods can be used to study this system and much theoretical work has been done so far (for a review see e.g. 6 ). The Hartree-Fock method gives a simple analytical solution, and though it totally ignores the effect of correlations, it provides a useful insight on the structure of the single particle levels. Another interesting approach consists in applying a unitary transformation U 7,8 giving, to leading order in the spin-orbit strength, a transformed Hamiltonian H = U -1 HU whose eigenstates are also spin and angular momentum eigenstates. Though approximate, this method allows the use of standard Quantum Monte Carlo techniques 9 . | cond-mat.str-el |
( Here in order to compare the analytical results with the numerical one, the confining potential along the zdirection V c (z) is taken as Vc (z) = 0 -L z /2 < z < L z /2 ∞ otherwise,(2) where L z is the well width of the quantum well. The asymmetrical potential, which stems from a build-in electric field F via the gate voltage or δ-doping is Va (z) = eF z, which breaks the inversion symmetry and lifts the spin doublet degeneracy. Let Ŝ be the generalized spin operator of a hole state, and Ŝz be the z-component of Ŝ, the isotropic Luttinger-Kohn Hamiltonian ĤL in the |S, S z representation (four basis kets written in the sequence of {| 3 2 , | 1 2 , | -1 2 , | - 3 2 }) is expressed as ĤL = P R T 0 R † Q 0 T T † 0 Q -R 0 T † -R † P ,(3) with P = 2 2m 0 [(γ 1 + γ 2 )k 2 + (γ 1 -2γ 2 )k 2 z ],(4) Q = 2 2m 0 [(γ 1 -γ 2 )k 2 + (γ 1 + 2γ 2 )k 2 z ],(5)R = - 2 √ 3γ 2 m 0 k -k z ,(6)T = - 2 √ 3γ 2 2m 0 k 2 -,(7) where γ 1 , γ 2 is the Luttinger parameters, m 0 is the free electron mass, the in-plane wave vector k = (k x , k y ), denoted in the polar coordinate as k ≡ (k, θ), k ± ≡ k x ±ik y and k z = -i∂/∂z. The other terms, such as anisotropic term, C terms or hole Rashba term, 35,36,37 have only negligible effects and are omitted in our calculation. Correspondingly, the x-, y-, z-component of the "spin"-3 2 operator respectively reads Ŝx = 1 2 0 √ 3 0 0 √ 3 0 2 0 0 2 0 √ 3 0 0 √ 3 0 ,(8) Ŝy = i 2 0 - √ 3 0 0 √ 3 0 -2 0 0 2 0 - √ 3 0 0 √ 3 0 ,(9) Ŝz = 1 2 3 0 0 0 0 1 0 0 0 0 -1 0 0 0 0 -3 . (10) We stress here again that the "spin" of the 3 2 spinor is actually its total angular momentum, which is a linear combination of spin and orbit angular momentum of a valence band electron. | cond-mat.mes-hall |
This yields the U (1) holonomy U (b) = e i(K•b)τ3 = e i 2π 3 (b1-b2)τ3 ,(2) where b 1 and b 2 are the integer components of the Burgers vector b in the lattice basis (see Fig. 2). The dislocations thus separate into three equivalence classes, labeled by d ∈ {0, 1 3 , -1 3 }, with 3d ≡ (b 1 -b 2 ) mod 3, where the period of 3 follows from the periodicity of the Fermi states (see Fig. 2). Different from the case of disclinations, 6,7 this is independent from the A/B sublatitce pseudospin quantum number since translations carry no information on the structure inside the unit cell. Instead, this phase does depend on the valley quantum number in a simple way: the absolute magnitude is the same and the phases in the two valleys just differ by a minus sign. | cond-mat.mes-hall |