ISSP
Activity Report
2002
Contents

Pages

Preface1

Research Highlights

2 - 28Highlights of Joint Research

29 - 44

ISSP Workshop

45 - 47

ISSP Activity Report 2002

PREFACE- Activity Report -The ISSP was established in 1957 based on the recommendation of the National Council of Japan to
set up a new institute for condensed matter physics. The idea was to establish a top class research
institute for experimental and theoretical studies of condensed matter and to promote cooperative
research in the field throughout the country. Soon after the inauguration, the name of ISSP became
recognized world-wide in the community of condensed matter physics.
Since then there have been two major changes in the history of the institute. The first was the reorganization of the institute in the mid-eighties. The system of divisions consisting of several groups
was introduced and the project of the “Physics in extreme conditions’’, which includes Ultra-High
Magnetic Fields, Laser Physics, Surface Physics, Ultra-Low Temperatures, and Very High Pressures,
was started. The second major change was the relocation of the ISSP to the new Kashiwa campus
which was completed in 2000.
To make the status of the research activities public, new forms were introduced last year: Activity
report A for research mostly by the staff of the ISSP and Activity report B for joint research conducted by the ISSP. This is the Activity Report for the fiscal year 2002 in the new style of a single volume but keeping the two sections A and B.
We are happy to receive any comments on the Report for the possible improvement of the research
activities of the ISSP.

August 31, 2003

Kazuo Ueda
Director
Institute for Solid State Physics
The University of Tokyo

ISSP Activity Report 20021

Section A
Research Highlights
Spin Superstructures in the
1/8-Magnetization Plateau Phase of the2D Orthogonal Dimer Spin System
SrCu2(BO3)2
M. Takigawa and Y. Ueda
The discovery of a spin-gap and magnetization plateaus[1] in the quasi two-dimensional orthogonal dimer spin system SrCu2(BO3)2 stimulated vast amount of experimental
and theoretical activities. The crystal structure realizes a
stack of the two-dimensional Shastry-Sutherland spinmodel, for which the direct product of dimer singlets is theexact ground state. The triplet excitations are known to havean extremely small dispersion width. A striking property of
SrCu2(BO3)2 is the plateaus in the magnetization vs. field
curve at fractional values (1/8, 1/4 and 1/3) of the fully saturated magnetization. It has been argued that the excitedtriplets with small kinetic energy localize due to repulsive
interactions to form superlattices in the plateau phases, i.e.,the Mott insulating state or Wigner crystal of magnetic excitations are realized. Since the nuclei on inequivalent Cusites in a superstructure should have different magnetic
hyperfine fields, which can be observed by NMR, we have
performed 65Cu and 63Cu NMR measurements at low temperature (35 mK) in the magnetic field up to 28 T using thefacilities in the Grenoble High Magnetic Field Laboratory[2]. The field range covers the first 1/8 plateau state, which
occurs for 27 – 28.5 T for the field applied perpendicular to
the 2D layers.
Figure 1 shows the Cu NMR spectra at 26 T (inset) andat 27.6 T (main panel). Generally, one type of Cu-sites witha specific value of hyperfine field gives six NMR lines;
three lines split by electric quadrupolar interaction for eachof two isotopes 65Cu and 63Cu. The spectrum at 26 T canindeed be represented as superposition of such six lines,
indicating that the magnetization is largely uniform. At
27.6 T near the middle of the 1/8 plateau, we observed completely different spectrum with large number of sharp peaks
distributed over a wide frequency range, providing clear evidence for a magnetic superstructure breaking the translational symmetry.
We found that any satisfactory fitting of the spectrum
requires at least 11 Cu sites with different hyperfine fields.

Fig. 2. Right part: Magnetization profile obtained by exact diagonalization of the Shastry-Sutherland model. Red (blue) circles indicate
positive (negative) 〈Sz〉. The circle size represents the magnitude of〈Sz〉. Left part: Histogram of the hyperfine field distribution. The middle panel (red) shows results of the fitting of the NMR spectrum (Fig.
2, red line). Long (short) lines indicate that the population of the site
is 1/8 (1/16). The top panel is obtained from the calculated profile(right part) assuming only the on-site hyperfine coupling. The bottom
panel shows the result when the transferred hyperfine couplings,
denoted as B and C in the right part, are adjusted to best reproduce
the experimental results.

An example is shown by the red line in Fig. 2. In particular,
the six sharp lines in the red zone (105–165 MHz) are
ascribed to the sites with the largest magnetization〈Sz〉=0.30. Likewise the spectrum in the yellow zone (165235 MHz) represents another site with 〈Sz〉=0.20. The rest of
the spectrum comes from the sites with smaller hyperfine
fields. The distribution of the hyperfine field obtained bythis fitting is shown in Fig. 2 (middle panel in the left part).
It should be noted that some sites have positive hyperfine
fields, i.e., some spins are polarized opposite to the external
field. Thus the magnetization oscillates within the unit cellof the superstructure.
By symmetry consideration, we can conclude that theunit cell of the superstructure has the rhomboid shape shownin Fig. 2. The number of inequivalent Cu sites for the rhomboid cell is eight for a single layer, but it may increase up to
16 depending on the pattern of stacking layers along the caxis. We found that other types of unit cell cannot produce
as many as 11 sites.
The magnetization profile calculated by exact diagonalization of the Shastry-Sutherland model on a 16-spin clusterof the rhomboid cell is shown in Fig. 2. One can clearly recognize a building unit consisting of one strongly polarized
dimmer surrounded by decaying oscillation of magnetization. This resembles the Friedel oscillation near impurities inmetals. The distribution of the hyperfine fields derived fromthe calculatied magnetization profile captures the essential
feature of the experimental results, indicating that such amagnetic superstructure is indeed realized in the 1/8 plateau
phase.
References[1] H. Kageyama et al., Phys. Rev. Lett. 82, 3168 (1999).
[2] K. Kodama et al., Science 298, 395 (2002).

Fig. 1. Cu NMR spectra at 35 mK. The main panel (inset) shows the
spectrum in the field of 27.6 (26) T. The red line is the fit with 11
distinct hyperfine fields. The arrows in the main panel show the resonance frequencies for zero hyperfine field. The intense peaks for 371– 379 MHz (at 126 MHz) is due to 11B (10B) nuclei.

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ISSP Activity Report 2002

Authors
K. Kodama, M. Takigawa, M. Horvatica, C. Berthiera,b, H. Kageyama,
Y. Ueda, S. Miyaharab,c, F. Beccac, and F. MilacaGrenoble High magnetic Field Labrqatory, CNRS-and MPI-KFK,
bLaboratorie de Spectrométrie, Université Joseph Fourier,cInstitut de Physique Théorique, Université de LausanneNovel Superstructures in the Metal
Ordered Manganite YBaMn2O6
Y. Ueda and H. Yoshizawa
The physics of manganese perovskites, A3+1-xA'2+xMnO3,
is full of surprises. Emergence of various interesting phenomena such as colossal magnetoresistance is largely due tothe versatility of the A site, which forms a wide range ofsolid solutions between cations with different valences andionic sizes. This advantage provides us with an opportunity
to systematically tune the doping level as well as the tolerance factor. However, such a partial substitution in turncauses a structural disorder, leading to quantitative discrepancy between experiment and theory or even masking some
intrinsic features which should inhibit in a 'clean' system.
We have successfully obtained a new manganite perovskite free from a structural distortion, that is, YBaMn2O6
with the A site ordering in layers along the c axis. Amongprominent features observed in this material are successive
phase transitions containing four phases and an extremely
high metal-insulator transition temperature (480 K). In orderto obtain insight into the microscopic picture of the paramagnetic insulating (PI) phase below 480 K, we performed
transmission electron microscopy (TEM) using JEOL
EM2010 operated at 200 kV. At room temperature, we
observed a superreflection represented as q1=(1/4, 1/4, 0)p,
see Fig. 1(b). This can be explained in terms of acharge/orbital ordering (COO) of the CE type, which has
been mostly observed for the disordered systems with x
being close to 0.5. However, the [100]p-zone electron diffractions shown in Fig. 1(a) further revealed another modulation, q2=(0, 0, 1/4)p. Together with the result of Rietveld
refinement of powder neutron/x-ray diffractions, we haveconcluded that the superstructure consists of planes of the
CE type built up according to the sequence [ααββ...], seeFig. 2 (a). This superstructure, which has been observed forFig. 3. Neutron diffraction of YBaMn2O6 at 250 K and 20 K, where
the peaks are indexed according to the √2ap × √2bp × 2cp unit cell.

the first time, should be strongly correlated to the Y/Ba
order, since this can make 'orbital' interactions alternated,leading to the quadruplicate periodicity.
Figure 3 compares powder neutron diffraction patterns at
250 K and 20 K, obtained using the Kinken powder diffractometer, HERMES, at the Institute for Material Research,
Tohoku University, installed at the JRR-3M reactor in
JAERI, Tokai. The magnetic peaks at 20 K in the antiferromagnetic insulating (AFI) phase are characterized by theplanes of the CE type with the 4-fold periodicity along the caxis. It is natural to consider again that interlayer magnetic
interactions become alternated, ferro- and antiferromagnetic,
as a result of the A site ordering. Interestingly enough, theq2 modulation in the PI phase disappears in the AI phase,indicating the COO pattern of either [αααα...] or [αβαβ...].
Transformation of the charge/orbital ordering would be
mediated by the spin ordering. In other words, there is a
close interplay between spins and orbitals.
References[1] T. Nakajima et al., J. Phys. Chem. Solid 63, 913 (2002).
[2] H. Kageyama et al., J. Phys. Soc. Jpn 72, 241 (2003).
Authors
H. Kageyama, T. Nakajima, M. Ichihara, Y. Ueda, H. Yoshizawa, K.
OhoyamaaaTohoku University.

Fig. 1. High-resolution electron diffraction patterns of YBaMn2O6 at
RT obtained along (a) the [100]p-zone and (b) [001]p -zone axes,
where the subscript p denotes the notation for the simple cubic perovskite.

Fig. 2. Proposed charge/orbital and spin ordering patterns in the PI
and AFI phases, respectively.

A Light-Emitting Diode Fabricatedfrom Horse-Heart Cytochrome c
H. Tajima and H. Akiyama
Cytochrome c is a relatively low-molecular-weight protein which functions as an electron carrier in the living systems [1]. It belongs to the class of iron-porphyrin-containing
proteins, such as hemoglobin, myoglobin, and catalase. The
crystal structures of cytochrome c obtained from several living systems are similar at least around the iron-porphyrin
which is called as heme c. Figure 1 shows the chemical
structure of heme c. In this unit, the Fe atom which may
exist in either the oxidized Fe(III) or reduced Fe(II) state, isaxially coordinated by histidine (His) and methionine (Met)groups.
We have fabricated ITO/cytochrome c/Al junction fromhorse-heart cytochrome c. This junction works as a light-

ISSP Activity Report 20023

References[1] Robert A. Scott and A. Grant Mauk ed., "Cytochrome c",
University Science Books, Sausalito, California, 1996.
[2] H. Tajima, S. Ikeda, M. Matsuda, N. Hanasaki, Ji-Won Oh, H.
Akiyama, Solid State Commun., 126, 579 (2003).
[3] W. A. Eaton, R. M. Hochstrasser, J. Chem. Phys., 46, 2533 (1967).
[4] P. M. Champion, R. Lange, J. Chem. Phys., 73, 5947 (1980).
Authors
H. Tajima, S. Ikeda, M. Matsuda, N. Hanasaki, Ji-Won Oh, H.
Akiyama

Local and Reversible Restructuring ona Ge(001) Surface Driven by STM
below 80 K
F. Komori and Y. Yoshimoto
Fig. 1. Molecular structure of heme c (iron porphyrin) in cytochromec. The abbreviations Met, His denote methionine, and histidine
respectively.

emitting diode [2]. Figure 2 shows the absorption spectrumof the cytochrome c film on a glass plate and the typicalelectro-luminescence (EL) spectrum of the diode.
The EL spectrum exhibits peaks around 530 nm and690 nm, and a weak shoulder around 410 nm. The 410 nmband corresponds to the Soret band, and the 530 nm band to
the Q band. The shape of the EL spectrum is completely different from either the absorption spectrum [3] or the photoluminescence (PL) spectra [4]. (The oxidized cytochrome cdoes not show any fluorescence, and the reduced
cytochrome c shows a very low yield of the Q band fluorescence when excited in the Soret or Q band.)
Most interesting result obtained in this study is theobservation of the strong and broad band around 690 nm in
the EL spectrum. This emission band consists of several
peaks. The corresponding band very weakly appears in theabsorption spectrum. Up to now, there is no report for theobservation of this band in the PL experiment. The absorption band around 690 nm is assigned to the charge-transfer
excitation between Fe in the porphyrin ring and the Met ligand (See Fig. 1) [3]. Therefore, the appearance of this bandin the EL spectrum suggests that the interaction between Fe
and Met ligand plays a crucial role in the electrical conduction in the cytochrome c film.

Fig. 2. The typical EL spectrum of the ITO/cytochrome c/Al junction
and the absorption spectrum of a cytochrome c film formed on a glass
plate.

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ISSP Activity Report 2002

Reversible control of local surface superstructures by
STM is expected on the ground state c(4×2) reconstructedsurface of Ge(001) as well as Si(001) because the energy
difference between the ground state and the p(2×2) structureis estimated to a few meV/dimer [1]. We have demonstrated
that the surface reconstruction on Ge(001) is locally andreversibly changed between c(4×2) and p(2×2) by controlling the sample bias voltage, Vb , of a scanning tunnelingmicroscope (STM) below 80 K. The c(4×2) structure isobserved with Vb<-0.7 V while the p(2×2) with Vb>+0.8 V.
Moreover, the both structures can be kept under scanning
the surface with |Vb|<0.6 V after fixing the surface structureof a local area by scanning there with a higher |Vb|. Figures 1
shows successive local modifications of the surface. Aftertaking the image shown in Fig. 1(a), the area A in the figure
was scanned with Vb=0.8 V to change from c(4×2) top(2×2). The result is shown in Fig. 1(b). The change along
the dimer row continued to the outside of the scanning area.
Then we scanned the area B in Fig. 1(b) to form anotherp(2×2) area as shown in Fig. 1(c). The reverse change fromp(2×2) to c(4×2) was done by scanning the area C in Fig.
1(c) with Vb=-0.7 V. In this case, we had to scan all of thedesired p(2×2) area because the change to c(4×2) was limited nearby the area of the scanning along the dimer row. The

Fig. 1. (a-d) Successive STM images of the same area (22.5 × 10nm2) of a Ge(001) surface at 80 K with Vb=-0.2V. The square areas"A" in (a) and "B" in (b) were scanned with Vb=0.8 V to change thelocal structure from c(4×2) to p(2×2) after taking each image. The
surface after scanning "A" is shown in (b), and that after scanning"B" is in (c). The rectangular area "C" in (c) was scanned with Vb=0.7 V to change it from p(2×2) to c(4×2). The surface after scanning
the area "C" is shown in (d). (e) A schematic showing the observed
structures depending on the bias voltage and the change direction.

Fig. 2. When the surface is positively biased under the STM tip, thenegatively charged upper-atoms of the dimers are electro-staticallypushed toward the subsurface by the tip, and the lower-atoms arepulled toward the tip. Consequently, the dimer buckling becomes
smaller. This makes the difference of the subsurface strain energy
between c(4×2) and p(2×2) structures smaller than that of the dipole
energy among buckled dimers. Thus for the positively biased surface,
the p(2×2) structure is favored by the gain of the dipole energy, andcan be the ground state.

annihilation of p(2×2) is shown in Fig. 1(d). The observed
structure is summarized in Fig.1(e). The bistability betweenp(2×2) and c(4×2) on Vb is attributed to the changes in twocompeting energies determining the structure; the electric
dipole energy among the Ge dimers, and that of the subsurface lattice strain. The upper atom of the Ge dimer is negatively charged, and the dipole energy for p(2×2) structure issmaller than that for c(4×2). On the other hand, the strainenergy for p(2×2) is larger than that for c(4×2). Both areaffected by the electric field between the tip and the surface
as schematically shown in Fig.2. For the positively biased
surface, the p(2×2) structure is favoured by the gain of thedipole energy, and can be the ground state. On the other
hand, under the negative Vb, the buckling of the dimers isincreased by the electric field, and the c(4×2) structure ismore favoured. The current dependence of the transitionindicates that an inelastic scattering process during the electron tunnelling induces the transition under the electric fieldby STM.

Fig. 1. Schematic diagram of the whole view (above) and sample
chamber (below) of the opposed-anvil type apparatus using sintereddiamond.

anvil type apparatus has much simple structure and has highpotential to extend pressures, as is clear from diamond-anvil
apparatus. In diamond anvil, however, uniform and stable
heating of the sample is extremely difficult, because of the
intrinsic nature of laser heating. Here we used an opposedanvil apparatus adopting sintered diamond as anvil material,and developed sample assembly to make uniform and stable
heating by resistance heating. Fig. 1 is the schematic diagram of the sample assembly. The heater material is a mixture of TiC and diamond powder hardened by adding small
amount of epoxy resin. All the small parts were prepared bya laser machining system, which enables us to make small
size machining with high precision. Heater material, sample
chamber, and electrodes thus prepared are stacked in thehole of the pressure transmitting medium made of a mixtureof MgO and diamond, and then compressed by two sintereddiamond anvils. By running an electric power of about 200
W to this assembly, temperature higher than 2000 K can beachieved. Fig.2 is an example of the sample recovered from

Reference[1] Y. Yoshimoto et al., Phys. Rev. B61, 1965 (2000).
Authors
Y. Takagi, Y. Yoshimoto, K. Nakatsuji and F. Komori

Development of an Opposed-Anvil Type
Apparatus for High Pressure-Temperature
Melting Experiments
T. Yagi
A high pressure-temperature apparatus capable for uniform melting experiments at very high pressures is developed. Use of double stage multi-anvil apparatus made it possible to perform well controlled high pressure-temperatureexperiments up to about 25 GPa. Although the pressure
range is further extended recently by the use of sintered diamond as anvil material, multi-anvil system has various difficulties when experiments above 40 GPa are tried. Opposed-

Fig. 2. SEM photograph of the cross section of the sample recoveredfrom high pressure-temperature experiment. Three photographs at the
top are the enlarged pictures of various small parts. Various featuresof crystal growth, formation of immiscible liquid, and reactions
among ZrO2, TiC, Mo, and Re can be seen.

ISSP Activity Report 2002

5

high pressure-temperature experiments, which show a mixture of TiC, ZrO2, Mo, and Re, melted at pressures above
30 GPa. This is a preliminary test and further extension of
the pressure range is expected but this figure show highpotential to make melting experiment at very high pressureusing this new apparatus. Further development to makequantitative analysis of pressure and temperature are in
progress by combining with a synchrotron radiation.
Authors
T. Yagi, H. Goto, D. Rubiea, and D. FrostaaUniversity of Bayreuth

Superconducting Gap Structure ofUnconventional Superconductors
Y. Matsuda
Nodes of the superconducting gap functions are often
important hallmarks for unconventional superconductors.
Despite the fact that the gap functions have immediate bearings on the pairing mechanism, we have for most of theunconventional superconductors been lacking detailed
knowledge on its structure, especially the direction and typeof nodes present. A recent series of experiments carried out
by our group have demonstrated that the low temperaturethermal transport measurements in magnetic field rotated invarious directions is a powerful tool for probing such nodal
structures. We proved the superconducting gap structures of
2D superconductors (1) spin triplet Sr2RuO4, (2) heavy
fermion CeCoIn5, (3) organic κ-(BEDT-TTF)2Cu(NCS)2,
the 3D superconductor (4) borocarbide YNi2B2C and (5) the
new heavy fermion PrOs4Sb12, by the low temperature thermal transport measurements. The gap functions of Sr2RuO4,
CeCoIn5 and κ-(BEDT-TTF)2Cu(NCS)2, have line nodes
which are shown in Figures below. We also obtained evidence for the existence of point nodes for YNi2B2C andPrOs4Sb12 shown in Figures below. For PrOs4Sb12 weobserved an unexpected multiple superconducting phases.

Fig. 1. Superconducting gap structure of unconventional superconductors, which are proved by the low temperature thermal transportmeasurements.

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ISSP Activity Report 2002

References[1] K. Izawa et al., Phys. Rev. Lett. 86, 2653 (2001).
[2] K. Izawa et al., Phys. Rev. Lett. 84, 057002 (2001).
[3] K. Izawa et al., Phys. Rev. Lett. 88, 027002 (2002).
[4] K. Izawa et al., Phys. Rev. Lett. 89, 137006 (2002).
[5] K. Izawa et al., Phys. Rev. Lett. 90, 117001 (2003).
Authors
Y. Matsuda and K. Izawa

Field-Induced Ordering in the HeavyElectron Superconductor PrOs4Sb12
T. Sakakibara
The cubic skutterudite compound PrOs4Sb12 has recently
been attracting much interest because it is the first Pr-based
heavy-electron superconductor that is most probably of
unconventional type [1,2]. In PrOs4Sb12, the Van Vleck
behavior of the magnetic susceptibility suggests that the
crystalline-electrical-field (CEF) ground state is either anon-magnetic doublet or a singlet [3]. The heavy electron
behavior has often been discussed in relation to a quadrupolar Kondo effect assuming the doublet ground state of
Pr3+[1,3]. Identification of the CEF ground state of this system is of importance for understanding the formation of theheavy-electron state in this system, but the situation seems
to be still controversial [3,4]. In order to gain further insight
into this issue, we examined the field-induced ordering in
PrOs4Sb12 [5] which has recently been revealed for H||[100]by the specific heat measurements [4]. Since this type ofphase diagram has never been observed in Ce-based heavy
fermion compounds, the origin is considered to be specificto the 4f electronic state of Pr3+.
Field-induced phase transitions can be seen most clearlyin the differential susceptibility data dM/dH, some examplesof which are shown in Fig. 1 for H||[110]. Here the transition to the ordered phase (A-phase) is marked by a jump in
dM/dH at ~50 kOe (thick arrows). A transition from the Aphase to a paramagnetic state can also be seen at ~110 kOe.
For this direction, we observed additional peaks in dM/dH at
H1 and H2 within the A-phase, possibly due to certainchanges in the ordered structure.

Fig. 1. Differential susceptibility of PrOs4Sb12 for H||[110], obtained
at T = 0.06 K~0.68 K. The field-induced ordering is marked by ajump in dM/dH as shown by thick arrows. Two distinct peaks areobserved in dM/dH at H1 and H2 within the ordered phase, suggestingsome changes in the ordered structure. The sharp structures below
Hc2 are due to the peak effect. Small oscillations seen above 8 T are
the de Haas-van Alphen effect.

Fig. 2. H-T phase diagram of PrOs4Sb12 with H||[100], [110] and[111] as indicated. Field-induced ordered phase (A-phase) isobserved for all the principal directions. Double superconducting
transition is observed at low field. No strong anisotropy is found forthe superconducting state.

We continued the measurements also for the other directions, and obtained the field-temperature phase diagrams asshown in Fig.2 [5]. It should be emphasized that the orderedphase is observed for all the principal directions. The transition temperature TA(H) is somewhat anisotropic, being lowest (highest) for H||[110] ([100]). We also confirmed that the
A-phase closes at high fields, except for H||[100] where theupper border was not observed because of the present experimental limitation. The field-induced ordering is considered
to result from a level crossing of the Zeeman split CEF
states as well as intersite multipole interactions. It can beshown that the presence of A-phase, irrespective of field
direction, is incompatible with the doublet ground statemodel, indicating that the CEF ground state of this system isa Γ1 singlet [5]. This implies that the CEF excited state(triplet) existing at ~6 K above the ground state is responsible for the heavy-electron formation. As for the superconducting state, we observed that the double transition at
Tc1=1.85 K and Tc2=1.75 K recently reported in the specificheat measurement [3] extends to lower temperatures down
to T≈0 [5]. PrOs4Sb12 exhibits rich phase diagrams. Further
experiments are in progress to understand this interesting
system.

Fig. 1. Field dependence of the electrical resistance of TPP[FeIII(Pc)(CN)2]2 at 20 K. The field is applied parallel to the a-axis. The
inset shows the crystal structure. Note that the resistance decreases
less than 1 % of the zero-field value.

posed of inorganic anions having local magnetic moments
and molecular donors affording π-conduction electrons. The
key to synthesize such compounds is the strength of the d-πinteraction (the interaction between local magnetic momentsand π-conduction electrons). In the case of the charge-transfer complexes above mentioned, however, this strength isdetermined by the crystal structure whose design is still
beyond control. One of the ways to solve this problem is tosynthesize charge transfer complexes by using the molecules, where the strong d-π interaction is self-contained. In
this context, we are studying a new series of charge-transfer
salts containing the axial-substituted phthalocyanine (Pc),[FeIII(Pc)(CN)2]-, where the strong d-π interaction is selfcontained. Interestingly, all the conducting salts
containing [FeIII(Pc)(CN)2]- (TPP[FeIII(Pc)(CN)2]2 [1, 2],[3],
PTMAx[FeIII(Pc)(CN)2]·PXX[FeIII(Pc)(CN)2]
III
III
y(MeCN) [4], TPP[Fe 0.3Co 0.7(Pc)(CN)2]2[5]) exhibit theanisotropic giant negative magnetoresistance.
Recently, we found the resistivity of TPP[FeIII(Pc)(CN)2]2
decreases down to less than 1 % of the zero-field value
under the magnetic field of 38 T. (See Figure 1.) Thisdecrease is not due to a phase transition.
Figure 2 shows the normalized resistivity of
TPP[FeIII(Pc)(CN)2]2 plotted versus magnetic field at vari-

References[1] E.D. Bauer et al., Phys. Rev. B 65, 100506(R) (2002).
[2] K. Izawa et al., Phys. Rev. Lett. 90, 117001 (2003).
[3] R. Vollmer et al., Phys. Rev. Lett. 90, 057001 (2003).
[4] Y. Aoki et al., J. Phys. Soc. Jpn. 71, 2098 (2002).
[5] T. Tayama et al., J. Phys. Soc. Jpn., 72, 1516 (2003).
Authors
T. Tayama, T. Sakakibara, H. Sugawaraa, Y. Aokia and H. Satoa
aDepartment of Physics, Tokyo Metropolitan University

Giant Negative Magnetoresistance in
TPP[FeIII(Pc)(CN)2]2 and the Related
Compounds
H. Tajima
A search for a new function based on the interplay
between local magnetic moments and conduction electrons
is a current topic in the field of molecular conductors. Mostof research works concern charge-transfer complexes com-

Fig. 2. Normalized resistivity of TPP[FeIII(Pc)(CN)2]2 plotted versus
field at various temperatures. Note that the negative magnetoresistance is observed even at 70 K.

ISSP Activity Report 2002

7

ous temperatures. As can be seen from this figure, the negative magnetoresistance is observed even at 70 K.
References[1] M. Matsuda, T. Naito, T. Inabe, N. Hanasaki, H. Tajima, T.
Otsuka, K. Awaga, B. Narymbetov, and H. Kobayashi, J. Mater.
Chem. 10, 631 (2000).
[2] N. Hanasaki, H. Tajima, M. Matsuda, T. Naito, T. Inabe, Phys.
Rev. B 62, 5839 (2000).
[3] N. Hanasaki, M. Matsuda, H. Tajima, T. Naito, T. Inabe, Synth.
Met., 133, 519 (2003).
[4] M. Matsuda, T. Naito, T. Inabe, N. Hanasaki, H. Tajima, J. Mater.
Chem. 11, 2493 (2001).
[5] H. Tajima, N. Hanasaki, M. Matsuda, F. Sakai, T. Naito, T. Inabe,
J. Solid State Chem., 168, 509 (2002).
Authors
N. Hanasaki, M. Matsuda, K. Masuda, H. Tajima, E. Ohmichi, T.
Naitoa, T. Inabea
aDivision of Chemistry, Graduate School of Science, Hokkaido
University

Metallic States in the Novel HydrogenBonded System:β"-(CnDT-EDO-TTF)2X(n = 5, 6, 7)
H. Mori
In molecular conductors, it is of a great importance tocontrol a molecular arrangement, which determines the electronic state of the charge transfer (CT) salt. However, theintermolecular interactions are generally small, so that to
obtain only the desired phase among the polymorphism isusually difficult.
In organic solids, the hydrogen bond is known to afford
relatively large intermolecular interactions [1], by use ofwhich a molecular arrangement would be preferentially constructed. Moreover, the strength and dimensionality of thehydrogen-bonded network would be controllable by theFig. 1. Molecular structures of CnDT-EDO-TTF.

Fig. 3. Temperature dependences of electrical resistivities forβ"-(CnDT-EDO-TTF)2AsF6. The resistivity of the n = 5 salt is proportional to ~T 2.4, suggesting the electron-correlated system. The
resistivity change from the n = 6 to the 5 salt indicates the chemicalpressure effect in substitution of the steric hindrance part of the molecule.

introduction of a steric hindrance part to the molecule, so as
to regulate the electronic state from an insulator, a superconductor, to a metal. Therefore, we synthesized a novel seriesof CnDT-EDO-TTF (1 – 4; n = 5 - 8 in Fig. 1) containing
cycloalkylene and ethylenedioxy groups, owing to function
as a hydrogen bond (CH··O) and a steric hindrance, and carried out the crystal structure analyses, the band calculations,and electrical conductivities of D2X (D = 1 - 4) to clarify theelectronic state of this system.
The crystal structure analyses of the prepared single
crystal, β"-22PF6, revealed the layered structure composedof the donor and the anion sheets. In the donor layer, theCH···O contacts (< 2.89 Å) between cycloalkylene and ethylendioxy groups are observed and the two-dimensionalhydrogen-bonded network spreads in the ac plane. (Fig. 2)
Consequently, the calculated overlaps along the molecular
0º, 30º, and 60º directions are large enough to construct a
pseudo-two dimensional Fermi surface. In addition, the analogues, β"-(CnDT-EDO-TTF)2X (n = 5, 6, 7, X = PF6,
AsF6), have the similar donor arrangement in use of hydrogen-bonded interactions.
In order to compare the electronic states ofβ"-(CnDT-EDO-TTF)2AsF6 (n = 5, 6, 7), the electricalresistivity measurements were carried out. As shown in Fig.
3, the resistivity of β"-12AsF6 is proportional to ~T2.4, suggesting the electron-correlated system. The typical exampleof the correlated system is κ-ET2X [2]. In the most pressurized condition, the resistivity behavior corresponds to that of
the n = 5 salt. With releasing the pressure, the convex resistivity curve like that of the n = 6 salt was observed.
Therefore, the chemical pressure effect could be realized inβ"-(CnDT-EDO-TTF)2X (n = 5, 6) in the two-dimensionalhydrogen-bonded network.
References[1] G. R. Desiraju, Acc. Chem. Res., 29, 441-449 (1996).
[2] H. Urayama-Mori, H. Yamochi, G. Saito, K. Nozawa, T. Sugano,
M. Kinoshita, S. Sato, K. Oshima, A. Kawamoto, and J. Tanaka,
Chem. Lett., 1988, 55; K. Kanoda, Physica C, 282-287, 299 (1997).

Fig. 2. Crystal structure of β"-(C6DT-EDO-TTF)2PF6. The intermolecular CH…O contacts (< 2.89 Å) are indicated by the dotted lines.
The hydrogen network spreads in the two-dimensional (2D) plane.
Consequently, the peudo-2D electronic state is constructed.

8

ISSP Activity Report 2002

Authors
H. Suzuki, M. Suto, K. Yamashita, T. Maejima, S. Kimura, H. Mori,
H. Moriyamaa, Y. Nishioa, and K. Kajitaa
aToho University

Theory of Nonequilibrium Kondo Effectin a Quantum Dot
K. UedaRecently it has been established that the Kondo effect isrelevant to the transport property through a quantum dot.
Clearly a new feature of the Kondo transport compared withthe usual Kondo effect of magnetic impurities is the nonequilibrium nature, since the current is measured with afinite bias voltage.
To treat the correlation effect, the second order perturbation theory with respect to the Coulomb interaction U in thedot [1] has been used as one possible method. According tothis theory the Kondo resonance is simply suppressed by afinite voltage and does not show any particular structure in
the nonequilibrium situation. However, it is not clear
whether the second-order perturbation theory works well in
the nonequilibrium conditions. On the other hand, the noncrossing approximation (NCA) has predicted the double
peak structure at the chemical potentials of both leads [2].
This result was obtained also by other approaches, the equation of motion [2], a real-time diagrammatic formulation [3]and scaling method [4]. Moreover the splitting of the Kondo
peak was successfully observed by introducing a potential
difference in the source lead [5]. These are favorable results
for the splitting of the Kondo resonance by a finite bias voltage.
It is still not clear how the effects of the double peakstructure, if it is true, appear in the differential conductance.
Moreover concerning the NCA it is well known that the analyticity is broken in the low temperature limit. Thus one cannot discuss the conductance in the unitary limit by the NCA.
In this situation it is clearly necessary to study the nonequilibrium Kondo effects by a better theoretical approach.
We have developed a theory of quantum transport
through a dot under finite voltage by using the perturbation
theory up to the fourth order of U based on the Keldysh formalism.
1) The density of states is shown in Fig.1 in units of 1/πΓ.
With a small voltage, the Kondo peak is suppressed but
still keeps a single peak with additional broadening. A
remarkable result of the theory is that the Kondo peak
splits into double peaks near the chemical potentials of thetwo leads when the bias voltage exceeds the Kondo temperature, kBTK.
2) In Fig.2 the differential conductance as a function of bias
voltage is shown in units of 2e2/h for various values ofU/Γ. It is seen that the new peak appears for large U,
U/Γ = 6 between the zero-bias peak and the broad peak at

Fig. 1. The density of states are shown in units of 1/πΓ for various
eV/Γ where U/Γ = 6.

Fig. 2. The differential conductance in the unit of 2e2/h is denoted
about various values of U/Γ. The zero-bias peaks reach to 2e2/h in theunitary limit for the symmetric case. The anomalous peak in
eV>kBTK is developed with increasing U/Γ in addition to the zero
bias peak.

around eV~U. The new peak appears when the bias voltage exceeds the Kondo energy.
3) It may be possible to speculate that the present results are
relevant to the 0.7 conductance anomaly. In experiment[6], it is clearly demonstrated that the crossover between
the region of the conductance of the unitary limit and the
0.8 plateau region takes place around the Kondo temperature. This result is consistent with the prediction of thepresent study.
References[1] S. Hershfield et al., Phys. Rev. Lett. 67, 3720 (1991); Phys. Rev.
B. 46, 7046 (1992)[2] Y. Meir, N. S. Wingreen and P. A. Lee, Phys. Rev. Lett. 70, 2601
(1993); N. S. Wingreen and Y. Meir, Phys. Rev. B. 49, 11040 (1994).
[3] J. König et al., Phys. Rev. B 54, 16820 (1996).
[4] A. Rosch et al., Phys. Rev. Lett. 87, 156802 (2001); condmat/0202404.
[5] S. De Franceschi et al., Phys. Rev. Lett. 89, 156801 (2002).
[6] S. M. Cronenwett et al., Phys. Rev. Lett. 88, 226805 (2002).
Authors
T. Fujii and K. Ueda

Thermodynamics and Correlation
Functions of Spin-1/2 One-Dimensional
Heisenberg Models
M. Takahashi
The spin-1/2 Heisenberg chain
is theone of the fundamental models which has been extensively
investigated in the study of the magnetism in low dimensions. The model is solvable by Bethe Ansatz method andlots of physical quantities have been calculated exactly overpast 70 years [1]. However, many new developments arestill being reported continuously. Among them we haveachieved two significant progresses recently.
Thermodynamics and High-Temperature Expansion
The thermodynamics of the spin-1/2 Heisenberg chain
has been studied traditionally through the Thermodynamic
Bethe Ansatz (TBA) equations and the Quantum Transfer
Matrix (QTM) method [1]. Both methods finally result in akind of non-linear integral equations among some auxiliary
functions, from which we can evaluate the free-energy ofthe system. Quite recently we have found one more different
integral equation, which determines the free-energy rather
directly [2]. The new integral equation consists of one
unknown function and is quite simple. Particularly we have

ISSP Activity Report 2002

9

Here ζ(s) is the Riemann zeta function. However, we haverecently succeeded in calculating the third neighbor correlation function [4]Fig. 1. Specific heat for the spin-1/2 anti-ferromagnetic Heisenberg
chain. [12,12] and [50,50] Padé are calculated from the high temperature expansion of the order (J/T)24 and (J/T)100, respectively.

found the equation is extremely powerful for the calculationof the high temperature expansion (HTE) of the free-energy
as well as other thermal quantities [3]. Actually by expanding the integral equation with respect to J/T, we could obtainthe HTE of the specific heat C(T) and the magnetic susceptibility χ(T) up to the order of (J/T)100. If we use the standard
linked-cluster algorithm, only the order of (J/T)24 is attainable. Thus our results will present the fundamental benchmarks for the application of the HTE, especially to the quantum critical phenomena. Fig.1 and Fig.2 plot the specificheat and the magnetic susceptibility of the spin-1/2 anti-ferromagnetic (J>0) Heisenberg chain, calculated from thenumerical solution of the integral equation. The HTEs arealso shown in the plots. Note that, as the plain HTE series
always diverge around T/J =0.55, we have applied the Padé
approximation to the HTE series. Particularly the [50,50]
Padé’s are obtained from our HTE results of the order(J/T)100. One can see our new [50,50] Padé approximants
provide the precise values for C(T) and χ(T) even in very
low temperature region.
Exact Calculation of Correlation Functions
Evaluation of the correlation functions for the interacting
systems are critically significant but very difficult problem
even for the exactly solvable models. As for the antiferromagnetic Heisenberg chain
only
the nearest-neighbor and the next nearest-neighbor correlation are known for a long time as [1]Fig. 2. Magnetic susceptibility for the spin-1/2 anti-ferromagnetic
Heisenberg chain. [12,12] and [50,50] Padé are calculated from thehigh temperature expansion of the order (J/T)24 and (J/T)100, respectively.

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ISSP Activity Report 2002

Our calculation is based on the multiple integral representation of the correlation functions. Although the integral
representation itself was obtained already in 1990’s by
Jimbo et al, its analytical evaluation has not been accomplished until quite recently. However, Boos and Korepin
could devise a general method to evaluate the multiple integrals recently. We have found their method is actually applicable to the calculation of arbitrary correlation functions of
the spin-1/2 Heisenberg chain. Furthermore we could generalize the method to the Heisenberg XXZ chain (anisotropic
extension of the Heisenberg chain) in [5].
References[1] M. Takahashi, Thermodynamics of One-Dimensional Solvable
Models (Cambridge University Press, Cambridge 1999).
[2] M. Takahashi, in Physics and Combinatrics, eds. A.K. Kirillov andN. Liskova, p299-304, (World Scientific, Singapore, 2001).
[3] M. Shiroishi and M. Takahashi, Phys. Rev. Lett. 89 117201(2002).
[4] K. Sakai, M. Shiroishi, Y. Nishiyama and M. Takahashi, Phys.
Rev. E. in press.
[5] G. Kato, M. Shiroishi, M. Takahashi and K. Sakai, J. Phys. A :Math. Gen. in press.
AuthorM. Takahashi

Interplay of Orbital, Spin and Lattice inTitanium Oxides
M. Imada
Magnetic orders are frequently controlled and determined from underlying orbital orders in strongly correlated
electron systems. Lattice distortions and orbital polarizations
are also frequently influenced by magnetic structures. These
entanglements and interplays open a possibility of controlling transport, optical and magnetic properties by various
external control parameters like magnetic fields, electricfields, light, pressure etc. From this motivation, rich magnetic and orbital phases, their phase transitions and interplays
have attracted much interest in strongly correlated electron
systems, particularly in transition metal (TM) oxides.
Cuprate superconductors and manganese perovskite compounds with colossal magnetoresistance belong to the TM
oxides with 3d eg bands at the Fermi level. Perovskite-type
Ti oxides RTiO3 with R being a rare-earth ion belong to thesame type but with 3d t2g bands at the Fermi level and show
a very different behavior. Particularly, while the above
mother materials are commonly typical Mott insulators,
RTiO3 shows markedly rich and complicated behaviors possibly due to the above interplay. The crystal structure of
RTiO3 is a pseudocubic perovskite with an orthorhombic
distortion (GdFeO3-type distortion) and exhibits a magnetic
phase transition as a function of the magnitude of this distortion. LaTiO3 has relatively small GdFeO3-type distortionand shows AFM(G)-type magnetic structure below 130 K.
The origin of this G-type antiferromagnetism has long beenFig. 1. (a) Displacements of the R ions along the b-axis due to the
GdFeO3-type distortion. In RO-plane 1, they shift in the positive
direction while in RO-plane 2, they shift in the negative direction. (b)In the case of site 1, the distances between the Ti ion and R ionslocated in the (1,1,-1)-directions decreases. As a result, the La ionsgenerate the crystal field which is similar to the trigonal crystal fieldwith [1,1,-1] axis. Displacements of the O ions along the c-axis which
are induced by the tilting of the octahedron are also presented. (c)Energy-level structure in the crystal field due to the R ions. The
threefold degeneracy of the cubic-t2g levels split into three nondegenerate levels.

puzzling and controversial, because naively the orbital
degeneracy leads to the ferromagnetic ground state. To fully
understand electron correlation effects in perovskite compounds and in more general, in d-electron systems, it is recognized that puzzling properties of Ti perovskites have to be
explained by a convincing theory. The Ti perovskite compound also offers touchstone materials for the understandingof orbital physics, which is a hot topic of the physics of electron correlation.
In our study it has turned out that a very simple mechanism of the R crystal field resulted from the displacement of
R ion caused by the well-known GdFeO3-type distortion
leads to a consistent picture for the experimental results. We
have examined the crystal field of La cations by considering
the experimentally observed distortion of the La ions. The
crystal field Hamiltonian is constructed by using the experimental position parameters. Based on this Hamiltonian, we
have shown that the distortion of the La ions caused by the
GdFeO3-type distortion generates the crystal field withlower symmetry, which is similar to the trigonal crystal
field. Then the threefold degenerate cubic-t2g levels split
into three isolated levels. The energies and the spinexchange constant calculated by the effective Hamiltonianshow that the lowest-level occupations in this crystal fieldstabilize the AFM(G) state (see Fig. 1). Since the perovskitetitanates attract an enormous amount of interest, a number ofexperiments have been done so far. However, there had beenno theory which can explain these experimental results in a
unified way. Now the following experimental results are
explained within our theory [1,2].
(1): Emergence of the antiferromagnetic(G) (AFM(G))
states.
We have shown that the AFM(G) state is realized with

quenched orbital degrees of freedom due to the crystal
field from R ions.
(2): Neutron scattering [3]Recent neutron scattering experiment showed the spinwave spectrum characterized by isotropic spin couplings
as well as a small spin gap for LaTiO3. These characteristics are well reproduced in our picture. The isotropic spin
coupling is caused by the isotropy of transfer amplitudes
in our theory.
(3): Behavior of the Néel temperature in the magnetic phase
diagram.
According to the experimentally obtained magnetic phase
diagram [4], the Néel temperature monotonically decreases as R goes from La, Pr, Nd to Sm with increasing
GdFeO3-type distortion. The calculation of the spin coupling based on our theory well reproduces this behavior.
(4): Resonant x-ray scattering.
Recent resonant x-ray scattering experiment showed theevidence of an orbital ordering in the AFM(G) compoundsof RTiO3 (R=La, Pr, Nd and Sm) and that the symmetryof the orbital ordering is the same as that in the FM compounds YTiO3 and GdTiO3 [5]. Our present calculation of
the orbital state based on our theory well explains thisresult.
(5): X-ray diffraction measurement.
We have shown that the crystal fields from La cations stabilize an orbital ordering in which (xy-yz-zx), (xy-yz-zx),(xy+yz+zx) and (xy+yz+zx) are approximately occupied atfour Ti sites in the orthorhombic unit cell. This orbital
ordering is expected to induce nearly trigonal distortionsof TiO6 octahedra. These distortions were actuallyobserved by recent x-ray diffraction measurements [6,7].
(6): NMRA NMR spectrum for LaTiO3 can be well reproduced if
we take an orbital ordering model predicted in our theory[8].
Through this study, it has clarified that in the perovskite
type transition-metal oxides, a crystal field from the R ionscaused by the GdFeO3-type structure lifts the 3d orbital
degeneracy. This mechanism of lifting the orbital degeneracy competes with the usual Jahn-Teller mechanism, andeventually plays an important role in controlling the orbitalspin structures. The finding of this mechanism is of fair andgeneral importance since the GdFeO3-type distortion is auniversal phenomenon, which is seen in a large number ofperovskite-type compounds. This mechanism may also play
important roles on the electronic structures in other perovskite compounds.
References[1] M. Mochizuki and M. Imada, J. Phys. Soc. Jpn. 70, 2872-2875(2001).
[2] M. Mochizuki and M. Imada, cond-mat/0301049[3] B. Keimer, D. Casa, A. Ivanov, J. W. Lynn, M. v. Zimmermann,
J. P. Hill, D. Gibbs, Y. Taguchi and Y. Tokura, Phys. Rev. Lett. 85,3946 (2000).
[4] T. Katsufuji, Y. Taguchi and Y. Tokura, Phys. Rev. B 56, 10145(1997).
[5] M. Kubota, H. Nakao, Y. Murakami, Y. Taguchi, Y. Tokura, M. v.
Zimmermann, J. P. Hill and D. Gibbs: unpublished.
[6] M. Cwik, et al., cond-mat/0302087.
[7] J. Hemberger, et al., cond-mat/0303070.
[8] M. Itoh, T. Kiyama et al., unpublished.
Authors
M. Mochizuki and M. Imada

ISSP Activity Report 2002

11

Berry Phase and Spin Quantum Hall
Effect in the Vortex State of Superfluid
3He in Two Dimension
M. Kohmoto
The Berry phase arises in quantum-mechanical systems
with an adiabatic change on a closed loop in a parameter
space [1]. In spite of the fact that it is a phase of the wave
function, it could be related to physical effects and, in somecases, has a connection with topological numbers. The quantum Hall effect in Bloch electron systems is described as anadiabatic charge transport whose process is closed in aparameter space. A Berry phase is generated and the quantized Hall conductivity is written by the Berry phase as wellas the Chern number [2].
Recently, it has been pointed out that quasiparticles inthe vortex lattice of dx2-y2-wave superconductors are in the
Bloch states [3]. The spin quantum Hall effect occurs and its
conductivity is written by a Chern number [4]. Then, one
can expect that the spin Hall conductivity in the vortex stateis written by a Berry phase, when the effect can be described
as an adiabatic spin transport of a closed process.
Here, we discuss Bloch quasiparticles in the vortex stateof p-wave superfluid 3He in two dimensions. To realize 2D,
we introduce a strong confine potential along the z axis to
avoid quasiparticle excitations along the z axis. The boundary effect introduced by the confine potential locks the relative angular momentum of all the Cooper pairs in the same
direction along the z axis. Then, we take the z component of
the angular momentum is 1 in the whole region. The direction of the d vector becomes parallel to the angular momentum because of the existence of the magnetic dipole interaction which couples spin and orbit.
It is well known that the rotating superfluid is direct
analogy of type-II superconductors, and actually the vortex
states in superfluid 3He are detected by the experiments [5].
Then, we consider superfluid in a container that rotatesaround z axis. The vortex is assumed to be form a square lattice (see Fig1.). Because of the periodicity of the vortex lattice, the energy eigenstates of the quasiparticle are the Blochstates and parameterized by the momentum in the Brillouinzone.
We also consider that the system is in a magnetic fieldwith a weak and homogeneous gradient. Such a field cannot
be introduced in superconductors due to the Meissner effect.
The magnetic field couples to spin through the Zeeman term

Fig. 1. A square vortex lattice. The blue dots denote the vortex coresin the lattice. The black tin arrows denote the lattice periodicity.
Quantum theory admits the only symmetries which commute eachother, so the wave functions of the quasiparticles are the Bloch states
for the periodicity denoted by the cyan arrows.

12

ISSP Activity Report 2002

and does not to orbital currents because of the neutrality ofthe superfluid. The magnetic field could be represented byan adiabatically changing vector potential which couples tothe spin current. The adiabatic process is defined on a closed
loop in the Brillouin zone, so it generates a Berry phase.
The spin Hall current flows in the process above. We
calculated the expectation value of the total spin current
directly by using the adiabatic approximation and obtain thespin Hall conductivity. The conductivity is represented by
the Chern number and quantized when the quasiparticle has
an excitation gap as that in the d-wave vortex state. We haveshown that the spin Hall conductivity is written by the Berryphase. The spin transfer per a cycle per the boundary of theunit cell is quantized and related to the Berry phase.
We would like to comment on the fact that, in our argument, the orbital part of the pairing symmetry is not crucial
as long as the quasiparticle spectrum in the vortex state has
an excitation gap. The spin part is crucial because the spinrotational symmetry is needed to obtain well-defined spincurrents. Several authors have made efforts to find a way tomeasure spin transport [5]. Some experimental techniques todetect spin transfer are highly desirable.
References[1] M.V.Berry, Proc. R. Soc. London. Ser. A 392, 45 (1984).
[2] M.Kohmoto, Ann. Phys. (N.Y.) 160, 355 (1985).
[3] M.Franz and Z.Tesanovic, Phys. Rev. Lett. 84, 554 (2000).
[4] O.V.Lounasmaa and E.Thuneberg, Proc. Natl. Acad. Sci. U.S.A.
96, 7760 (1999).
[5] See, for example, J.E.Hirsch, Phys. Rev. Lett. 83, 1834 (1999).
Principal Publication and Authors
J. Goryoa and M.Kohmoto, Phys. Rev. B 66, 174503 (2002).
aAoyama Gakuin University

Possibility of a Metallic Phase in the
Half-Filled Hubbard-Holstein Model
Y. TakadaFor the last few decades, the competition between thephonon-mediated electron-electron attraction and the local
Coulomb repulsion has been discussed in various contexts
of condensed matter physics. Initially the problem wasraised in relation to the quasi-one dimensional organic
charge-transfer salts. In recent years, it has also become an
important issue in studying the perovskite oxides such as
cuprates and manganites.
Theoretically the Hubbard-Holstein (HH) model serves
as a useful framework to investigate the competition. In
addition to the kinetic-energy term represented by t the nearest-neighbor electron hopping integral, it consists of two
interaction parts. One is the term describing the electron correlation effects through U the on-site Coulomb repulsion as
in the Hubbard model. The other is the term to couple anelectron with local phonons as in the Holstein model, leading to αω0 the polaron stabilization energy with α the nondimensional electron-phonon coupling constant and ω 0 the
optic phonon energy.
For this model, there have already been many previous
works using various methods. These studies have revealed
that the HH model can afford interesting phase diagrams dueto the competition among charge-density wave (CDW),
spin-density wave (SDW), and superconductivity as wechange the parameters involved in the system, i.e., U, α, t,ω 0, and n the electron filling. Roughly speaking, the CDW

Fig. 1. Phase diagram in the (α,U) plane for the half-filled HubbardHolstein model.

(SDW) nature manifests itself in the ground state at half filling (n=1), if the effective electron-electron interaction Ueff(~ U-2αω 0 ) is negative (positive).
However, the detail of the CDW-SDW transition is notknown well. In fact, the appearance of a superconducting
phase is suggested [1] in the exact-diagonalization study on
the two-site HH model. Thus we carefully consider the transition region at half-filling by approaching the HH modelfrom the antiadiabatic region (t« ω 0 ) where the variable displacement Lang-Firsov canonical transformation is useful.
In order to avoid any error that might enter through theapproximations involved in dealing with the Coulomb correlation term, we have investigated the one-dimensional system so that the exact solution of Lieb and Wu can be taken
advantage of. By determining the displacement variationally, we have obtained fairly accurate results for the propertiesof the ground state. [2]In Fig. 1, we show the obtained phase diagram in the(α,U) plane at t=0.2ω 0. The metallic phase characterized by
the local spin moment L 0~3/8 and a small effective mass ofFig. 2. Superconducting Tc as a function of the lattice constant a0 in
A3C60 and its dependence on the crystal structure in comparison withexperimental results.

polarons appears along the line of U= 2αω 0, namely, in the
CDW-SDW crossover region. This phase seems to continuously convert into the superconducting phase in the attractive Hubbard model (or the system with α =0 and U<0).
We do not believe that this behavior is specific to one
dimension (1D); a similar metallic behavior will be seen inhigher dimensions as well, as long as |Ueff | is less than thepolaron bandwidth in the crossover region. We may evenconjecture that such a metallic phase will be realized moreeasily in higher dimensions in which we can avoid unusual
phenomena specific to 1D like the Mott transition at U=0
irrespective of t in the 1D Hubbard model.
It is not known for sure whether this kind of a metallic
phase has already been observed in actual materials or not,but it has been pointed out [3,4] that physical properties
including superconductivity in the alkali-doped fullerenes
A3C60 can be reproduced in a unified way in terms of thispicture. For example, the superconducting transition temperature Tc in the HH model with suitably chosen parameters is
obtained as a function of the lattice constant a0 as shown in
Fig. 2.
References[1] Y. Takada, J. Phys. Soc. Jpn. 65, 1544 (1996).
[2] Y. Takada and A. Chatterjee, Phys. Rev. B 67, 081102(R) (2003).
[3] Y. Takada, J. Phys. Soc. Jpn. 65, 3134 (1996).
[4] Y. Takada and T. Hotta, Int. J. of Mod. Phys. B 12, 3042 (1998).
Authors
Y. Takada and A. Chatterjeea
aUniversity of Hyderabad.

Aharonov-Bohm-Type Oscillations in
Triangular Antidot Lattice
Y. IyeHigh mobility two-dimensional electron gas (2DEG) at
GaAs/AlGaAs heterointerface provides an ideal stage to
explore various quantum transport phenomena in controlledenvironment. Use of microfabrication techniques allows us
to prepare a variety of artificial structures in mesoscopic
scale. Antidot lattice is one of such artificial structures inwhich high potential regions (andidot) for the 2DEGs are
periodically arranged to constitute a lateral superlattice asshown in Fig.1.
The right panel of Fig.1 shows a potential landscape anda scanning electron micrograph of a triangular antidot lattice. For this sample, the lattice period is a=960 nm and thediameter of the antidot is d=576 nm.

Fig.1. Left: Potential landscape of a triangular antidot lattice. Right:
Electron micrograph of a sample used in this study.

ISSP Activity Report 2002

13

What Causes Quantum Decoherence in
Mesoscopic Circuits?
S. KatsumotoFig. 2. Main Panel: Magnetoresistance of a triangular antidot lattice.
Top panel: Oscillatory part of the resistance near zero magnetic field.
The AAS(h/2e) and AB-type (h/e) oscillations are identified. Bottom
panel: oscillatory part of resistance at around the peak between the
ν=3 and 2.

The main panel of Fig.2 shows a magnetoresistance traceof the triangular antidot lattice sample measured at 30 mK.
Three types of oscillations have been observed. At low magnetic field, the so-called Altshuler-Aronov-Spivak (AAS)
oscillation with h/2e period and the Aharonov-Bohm (AB)type oscillation with h/e period are observed on top of alarge negative magnetoresistance. The top panel shows theoscillatory part of resistance in this region. The periods ofthese oscillations agree with those expected from to the unit
cell area.
In the quantum Hall regime at higher magnetic field,
another series of AB-type oscillations with periods longer
than the low-field counterpart are observed as shown on thebottom panel of Fig.2. The periods of the high field AB-type
oscillations reflect the area of edge state encircling eachantidot.
The AB-type oscillation is different in nature from theordinary AB oscillation in a single ring, which is known todiminish upon ensemble averaging. The AB-type oscillationin antidot lattice observed in a macroscopic sample muchlarger than the decoherence length is attributed to oscillatory
fine structures of the density of states, which arises fromunstable periodic orbits inherent to the chaotic system.
The temperature dependence of the amplitude of oscillations reveals characteristic decoherence mechanisms, which
are different among the these oscillatory phenomena andfrom those found in the single ring case.
Authors
M.Ueki, T.Terada, A.Endo, S.Katsumoto, Y.Iye

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ISSP Activity Report 2002

Quantum decoherence, which separates quantum
mechanics and classical one, has been a most important
issue in physics since the establishment of quantum mechanics. Quantum transport in mesoscopic systems is one of thebest laboratories of the decoherence. Here we present a possible mechanism of decoherence, i.e., spin-states in quantum
dots.
A quantum dot (QD) and an Aharonov-Bohm (AB) ring
are two extremes of quantum devices. The former emphasizes the particle nature of electrons through the single-electron charging effect, while the latter highlights them aswaves through the interference. We showed that in a QDAB ring composite system such as shown in Fig.1(a), bothcharacteristics of electrons emerge as the Fano effect [1].
Then does no decoherence occur in such composite systems? A possible mechanism is illustrated in Fig.1(c). Whenthe top electronic state is half-occupied with an electron, saywith up-spin, an electron with down spin can come into thedot. But the outgoing electron can be either with up-spin or
down-spin. If it is with up-spin, this cannot interfere withthe other part traversed through the “reference arm” part of
the ring, thus results in decoherence.
If this mechanism emerges, the coherence of the total
system should be largely affected by the number of electrons. That is, when the number is even, the top level shouldbe empty in the simplest model and there is no decoherence,
while half of traversing electrons cannot interfere when thenumber is odd. The number of electrons in a quantum dotchanges by one just at the peaks of Coulomb oscillation. The
amplitude of AB oscillation is a good measure of quantum
coherence. Hence the above inference leads to the predictionthat the AB amplitude should be largely changed when thegate voltage crosses peaks of Coulomb oscillation.
Figure 2 shows a result of measurement at a Coulomb
peak. As shown in the lower panel, the AB amplitude, thus
the quantum coherence is larger at the left side of thisCoulomb peak. Steep decrease and “lapse” of AB phase at

Fig. 1. (a) Schematic diagram of the sample circuit. (b) When the top"spin-pair" state is empty, an electron tunnels through the dot without
spin-flip. (c) When the top state is half-occupied, spin-flip process
which causes decoherence is possible.

Fig. 1. (a) A square grid pattern of 7 nm period on a flat nitrogenadsorbed Cu(001) surface. Clean Cu surface is imaged bright andc(2×2)N surface dark. Nitrogen density in this image is 1/4 of Cu surface atoms density on average. Sides of the square c(2×2) patches arealong 〈100〉 directions. (b) One dimensional array of the squarec(2×2)N patches on a N-adsorbed Cu(001) surface vicinal to [010]
direction. Width of the terrace is quantized by the period of the grid
pattern. Formation of these patterns is attributed to strain-relief on
clean Cu surface and monoatomic steps.

Fig. 2. Upper panel: 3D plot of conductance versus a plane of gate
voltage and magnetic field across a Coulomb peak. Small oscillation
against the magnetic field is AB oscillation. Lower panel: Extracted
AB component. Apparently the AB amplitude is larger in the left sideof the Coulomb peak. Phase lapse of π is also seen at the peak.

the peak is due to the two-terminal nature of the device. Thisasymmetry is reduced by comparatively high (-2 T) magnetic field due to the Zeeman splitting.
These features agree with the theoretical prediction andwe believe the observed asymmetry is due to the above
described decoherence mechanism. What happens then,when such spin-flip scattering makes up coherent Kondo
states? That would be a most important issue in our
research.

the nitrogen ion irradiation on the surface at room temperature. This pattern formation is mainly attributed to the relaxation of the elastic strain due to the lattice expansion on thenitrogen-adsorbed surface. Here, the clean Cu surface andthe monoatomic steps act as strain relievers although therole of the step is secondary for the formation of the grid
pattern.
We have found that step density and direction drastically
influence the formation of one-dimensional patterns on
Cu(001)-c(2×2)N vicinal surfaces [2]. Figure 1(b) shows
one-dimensional array of the N-adsorbed square patches on
the surface vicinal to the [100] direction. In this case, thewidth of the terrace is fixed to the period of the grid pattern.
On the surface vicinal to the [-110] direction, on the other
hand, a labyrinth pattern is formed as shown in Fig. 2(a).
Here, the lines of clean Cu surface are aligned along theReference[1] K. Kobayashi, H. Aikawa, S. Katsumoto and Y. Iye, Phys. Rev.
Lett. 88, 256806 (2002).
Authors
K. Kobayashi, H. Aikawa, A. Sano, S. Katsumoto and Y. Iye

Novel One-Dimensional Nanopatterns on×2)N Surfaces
Vicinal Cu(001)-c(2×F. KomoriSelf-organization of adsorbate-induced domains on solid
surfaces has been one of the important subjects of nanoscalescience. It has been shown that one- and two-dimensional
nanopatterns can be made on surfaces by adjusting the coverage of adsorbates and adsorption temperature. On a nitrogen-covered Cu(001) surface, there forms a grid pattern consisting of c(2×2)N square patches of 5×5 nm2 and 2 nm
wide clean Cu surface at half monoatomic layer coverage[1] as shown in Fig. 1(a). The surface was prepared byannealing a Cu(001) crystal to about 670 K in 5 min afterFig. 2. (a) A labyrinth pattern formed on a nitrogen-adsorbed Cu(001)surface vicinal to [-110]direction. The lines of clean Cu surface of 12 nm wide are along the 〈110〉 directions, and the size and shape of
the N-covered domains are inhomogeneous. (b) A stripe pattern on anitrogen-adsorbed Cu(001) surface vicinal to [-110] direction. Thispattern was obtained by annealing the surface in Fig. 2(a). The widthof N- adsorbed stripe domain parallel to the [-110] direction is smaller than that parallel to the [110] direction. Small areas with the squarec(2×2)N patches coexist.

ISSP Activity Report 200215

〈110〉 direction. The size and shape of the nitrogen-covered
domains are inhomogeneous on the surface. The image in
Fig. 2(b) shows the result of subsequent annealing of thesample in Fig. 2(a) for 5 min. The stripe domains parallel tothe [110] direction occupy three times larger area than thedomains with the stripe parallel to the [-110] direction, andsmall areas with the square patches coexist. We notice that
there are short strips of clean Cu surface along 〈100〉 directions at the corners of two long clean Cu lines along 〈110〉directions.
Main characteristics of the stripe patterns can be
explained by assuming an anisotropic expansion of the lattice at c(2×2)N surface; the lattice constant increases only
along the direction perpendicular to the clean Cu lines whilethat along the Cu line remains the same. No strain reliever isnecessary in the direction of the Cu lines while both the stepedges and the clean Cu surface are the relievers in the perpendicular direction. The observed short strips of clean Cu
surface end at the centers of the c(2×2) stripe domain, andnever connect two clean Cu lines. In the present model, theshort strips are formed only highly compressed surface at
the corners consistently with the general formation mechanism of the clean Cu lines.
References[1] F.M. Leibsle, C.F.J. Flipse and A.W. Robinson, Phys. Rev. B 47,
15865 (1993).
[2] S. Ohno, K. Yagyuu, K. Nakatsuji and F. Komori, Jpn. J. Appl.
Phys. 41, L1243 (2002).
Authors
S. Ohno, M. Yamada, K. Yagyuu, K. Nakatsuji and F. Komori

Site-Specific Reactions of the Buckled×2) at Low Temperature
Dimer on Si(100)c(4×J. Yoshinobu
The Si(100) surface is reconstructed to form a (2×1)
dimer row structure. Many experimental and theoreticalstudies have supported the buckled (asymmetric) dimer withthe c(4×2) phase as a ground state. One of the fundamental
interests on Si(100) is how the spatial localization of electron density of surface atoms controls the reactivity towards
different molecules. In the buckled dimer on Si(100)c(4×2),
a partial charge transfer occurs from a down dimer atom(Sd) to an up dimer atom (Su) and thus the Su and Sd atoms
become partially negatively and positively charged, respectively. Hence, the reactivity of the Su and Sd atoms is
expected to be highly selective towards Lewis-acid andLewis-base molecules. We studied the adsorption oftrimethylamine TMA and water molecules on Si(100)c(4×2)
using low temperature STM.
×2) at 80 K
TMA on Si(100)c(4×Trimethylamine (TMA) being a typical Lewis base molecule has drown much attention to study the fundamental of
Lewis base-silicon surface interaction. In spite of several
studies on this system, there is no direct experimental evidence about the site-specific adsorption of adsorbed TMA
on Si(100).
In Figures 1, many bright protrusions are observed for
the TMA adsorption on Si(100)c(4×2) at 80 K. The protrusions are randomly distributed on the surface. All the brightprotrusions are located at either side of the center of dimer

16

ISSP Activity Report 2002

Fig. 1. Occupied state STM images of the Si(100)c(4×2) surface
exposed to TMA at 80 K. V(sample) = -2.2 V, I = 0.1 nA. The
schematic adsorption model is also shown.

row and most of the protrusions are triangle-shaped. One of
the three corners of triangle shape protrusion is pointed
towards the dimer bond direction, and the center of the triangle-shaped protrusion is located on the down atom. We haveascribed the triangle shaped protrusion due to the three
methyl groups in the adsorbed TMA. Thus, the location andshape of protrusion clearly indicates that the molecularadsorption of TMA is purely site-specific on the down dimer
atom.
Water on Si(100) : C-type defectDefects on the surface play a crucial role not only inchemical reaction, thin film growth but also electronic properties at the surface and interface. One of the long-standing
unsettled problems in surface science is the origin and atomic structure of C-defects commonly observed on the Si(100)surface. The number of C-defects was increased by thewater exposure on Si(100) at both room and low temperature.
At 80 K, the C-defect appears as an asymmetric teardrop
shaped protrusion in the high-resolution unoccupied state
image as shown in Fig. 2. The protrusions are marked by U,
U′, D, and D′ depending on its position and direction of theasymmetry. The U and U′ (D and D′) are the mirror imageof each other (enantiomorphic) where the mirror plane isperpendicular to the dimer row. The U and D (U′ and D′) arealso enantiomorphic where the mirror plane is parallel to thedimer row. Different orientation of C-defects is also
observed in the occupied state image at 80 K. By virtue of

Fig. 2. Unoccupied state STM image of Si(100)c(4×2) at 80 K.
V(sample) = 1.0 V, I = 0.1 nA The asymmetric teardrop shaped protrusions are C-defects. The protrusions are marked by U, U′, D andD′ depending on its position and the direction of asymmetry.

the alternate buckled dimer arrangement on Si(100)c(4×2) at
low temperature (<200 K), the registry of C-defect is clearly
identified, i.e. the ‘head’ and ‘tail’ of the asymmetric protrusion lie on down dimer atom and the up dimer atom of theadjacent dimer, respectively. The results unambiguously
indicate that the C-defect is caused by the dissociative
adsorption of single water molecule on the same side of two
adjacent dimers.
Authors
M.Z. Hossain, Y. Yamashita, K. Mukai and J. Yoshinobu

×1 Surface
Dimer Buckling of the Si(001)2×
Observed by Low-Temperature Scanning
Tunneling Microscope
Y. Hasegawa
On the Si(001)2×1 surface, a dimer structure is formed
for a reduction of the dangling bond. The unique feature
which leads the surface to show a variety of phases is buckling of the dimer: the two dimer-composing atoms are located up and down. According to the generally accepted scenario on a temperature dependence of the structure, anorder-disorder phase transition occurs around 200 K; below
the temperature the buckled dimers are arranged with thec(4×2) ordering, and above the temperature the thermally
activated dimer flip-flop motion breaks the symmetry to
exhibit the 2×1 ordering. Recently, Kondo et al. [1] andYokoyama and Takayanagi [2], however, reported STM
images taken below the transition temperature showing symmetric dimers. Their results are inconsistent with the above
scenario, casting a doubt on the accepted ground state structure.
It has been argued that uncertainty of sample temperature could be the origin of the inconsistent observations. To
avoid the uncertainty, we first measured the temperature
with a resistance sensor attached on a dummy sample, andFig. 1. STM images taken below 10K with a sample bias voltage of 2.2 V (a) and +1.6 V (b). The filled state image (a) apparently shows
symmetric dimers while the empty state image (b) shows buckled
dimers. In an area surrounded by a dashed green line in image (a)(zoomed in (c)) a faint trace of the c(4×2) structure; alternative arrayof dark spots between the dimer rows, are observed, indicating an
existence of the c(4×2) periodicity and the buckled dimers there.

confirmed that the sample temperature during our observation is below 10 K [4].
Figure 1 shows STM images taken at 10 K. In the filledstate image (Fig. 1(a)) most of the surface area is seemingly
covered with symmetric dimers. In fact, however, forinstance in the zoomed area shown in Fig. 1(c), a very faintpattern of alternative dark spots between the dimer rows isfound. The pattern, with a periodicity of the c(4×2) structure, must be a trace of the intrinsic c(4×2) structure overlapped with the symmetric-dimer appearance. We believethat this is an evidence of the dimer buckling at this temperature. In the empty state image, presented in Fig. 1(b),almost all area is filled with zigzag dimer rows, supporting
for an presence of buckled dimers.
Apparent symmetric dimer imaging is observed only at
low temperature (<10 K). At an elevated temperature only
buckled dimers were observed in both polarities of the bias
voltage. We found that transition temperature between thesymmetric and buckled dimer imaging in filled state images
is around 40 K for the sample we used. It has been known
that in silicon the carrier concentration drops dramaticallyaround 40 K. Since in the filled state STM observation holes
are inherently injected, we speculate that the apparent symmetric imaging is related with the reduced screening effectto the injected charges. As a result of the charging, potential
at the surface is locally lowered, presumably making possible for the electrons in the π* states to be probed in the condition of the filled state imaging.
References[1] Y. Kondo, T. Amakusa, M. Iwatsuki, and H. Tokumoto, Surf. Sci.
453, L318 (2000).
[2] T. Yokoyama and K. Takayanagi, Phys. Rev. B 61, R5078 (2000).
[3] S. B. Healy, C. Filippi, P. Kratzer, E. Penev, and M. Scheffer,Phys. Rev. Lett. 87, 016105 (2001).
[4] M. Ono, A. Kamoshida, N. Matsuura, E. Ishikawa, T. Eguchi, and
Y. Hasegawa, Phys. Rev. B 67, R201316 (2003).
Authors
M. Ono, A. Kamoshida, N. Matsuura, E. Ishikawa, T. Eguchi, and Y.
Hasegawa

ISSP Activity Report 2002

17

Tuning the In-Plane Lattice Constant ofan Oxide Substrate
M. Lippmaa
The properties of thin heteroepitaxial oxide films often
deviate significantly from the properties of bulk crystals
with the same composition. The most common reason forthis is the presence of lattice strain in thin films, caused by amismatch in lattice constants of the film and a single-crystal
substrate. Systematic studies of the effects of lattice strain
are complicated by the very limited choice of high-quality
substrates available for thin film growth. The only practicalway of generating small strains in oxide lattices is to use astrain-control buffer layer. Traditional buffer layer design
relies on the gradual relaxation of a lattice constant in a mismatched heteroepitaxial layer. Residual strain can then be
tuned by adjusting the buffer layer thickness. This method
has a number of drawbacks, mainly because the final strain
depends on a variety of process parameters, such as growthrate, temperature, etc.
In order to avoid such problems, we have developed abilayer buffer for tuning the in-plane lattice constant of a
substrate. The final in-plane lattice constant on the buffer
surface depends only on the composition, which is easy tocontrol and is not sensitive to precise processing parameters.
We start by growing a thin BaTiO3 layer on a SrTiO3
substrate at a relatively low temperature of 650 oC. The
layer thickness can be 10 nm or less. A much thicker layerof Ba1-xSrxTiO3 is then grown on top of the BaTiO3 layer at
the same temperature. The films grow in a layer-by-layer
mode and contain, as expected, a large number of dislocations (Fig. 1(a)). Annealing the film at up to 1350 oC for 30
minutes allows the defects to migrate in the crystal. As
shown in a post anneal TEM image in Fig. 1(b), all thedefects migrate either to the highest mismatch SrTiO3/
BaTiO3 interface, or to the film surface. No defects could befound in TEM images of the annealed film. The final lattice
constant of the buffer is determined only by the Sr content
of the Ba1-xSrxTiO3 layer and can be freely tuned in the 0.39
to 0.40 nm range, a span of 2.5%. The top part of the bufferis also fully relaxed and has a step-and-terrace surface struc-Fig. 2. Reciprocal space mapping of (a) a Ba0.5Sr0.5TiO3 film on a
BaTiO3 buffer layer, (b) a Ba0.5Sr0.5TiO3 film grown directly on a
SrTiO3 surface at 650 oC, and (c) partial relaxation of the same film
after high-temperature annealing.

ture.
The importance of inserting a thin highly mismatched
BaTiO3 layer between the main buffer and the substrate canbe seen in Fig. 2. The reciprocal space map in Fig. 2(a)clearly shows that the Ba1-xSrxTiO3 layer is relaxed and thatthe in-plane lattice constant of the thin BaTiO3 layer isdetermined by the top part of the bilayer buffer. If the
BaTiO3 layer is omitted and a Ba1-xSrxTiO3 film is grown
directly on a SrTiO3 surface, partial lattice constant relaxation occurs, as shown in Fig. 2(b). High-temperatureannealing does not improve the buffer quality (Fig. 2(c)).
The presence of a highly mismatched layer is thus crucial
for trapping strain-relieving dislocations at the substrateinterface.
Reference[1] K. Terai, M. Lippmaa, P. Ahmet, T. Chikyow, T. Fujii, H.
Koinuma, M. Kawasaki, Appl. Phys. Lett. 80, 4437 (2002).
Authors

K. Terai a, M. Lippmaa, T. Ohnishia Spring-8.

Novel Magnetotransport Phenomena in
Low-Dimensional Bloch Electron Systems
under Strong Magnetic Fields
T. Osada

Fig. 1. Moire images generated from high-resolution cross-sectionaltransmission electron microscope images, showing the locations ofdefects in BaTiO3 films grown on SrTiO3. (a) A film grown at
650 oC has many defects. (b) All defects migrate to the film/substrateinterface after annealing the sample at 1100 oC.

18

ISSP Activity Report 2002

(1) Harper Broadening under Megagauss Fields
The RF technique for transport measurement has beendeveloped under pulsed ultra-high magnetic fields generated
by the single-turn coil device (B<120 T) and the electromagnetic flux compression device (B<300 T). Using these technique and devices, we have successfully measured the magnetoresistance of a quasi-one-dimensional (Q1D) conductor
NbSe3 up to 300 T. In the low temperature phase of NbSe3,
there remains only one small closed Fermi surface as aresult of band reconstruction due to two independent CDW
transitions. This small Fermi pocket causes the Shubnikovde Haas (SdH) oscillations below ~50 T. In the magnetic
quantum limit above ~50 T, where only the lowest Landau
subband is occupied, magnetoresistance has a broad peakaround 65 T, and decreases monotonously above it, andfinally returns to the level of zero-field resistance (Fig. 1).
This anomalous behavior strongly suggests the occurrenceof the Harper broadening of the lowest Landau subband dueorganic conductor α-(BEDT-TTF)2KHg(SCN)4. Interlayer
current under strong electric fields is measured using pulse
technique to avoid sample heating. As seen in the Fig.2,
Lebed resonance peaks split to double peaks as the electricfield is increased. From the splitting, we could estimate theFermi velocity as 7.5×104 m/s.
Authors
T. Osada, E. Ohmichi, S. Ikeda, T. Inokuchi, K. Kobayashi, M.
Kobayakawa, A. Ogasawara, and E. Komatsu

Fig. 1. Magnetoresistance of NbSe3 measured by the RF transmission
technique under ultra high magnetic fields up to 300 T. The data
above 120 T were obtained using the electromagnetic flux compression system (inset), whereas those below 120 T were obtained using
the single-turn coil system. Magnetoresistance shows monotonous
decrease above 70 T, which originates from the Harper broadening of
the lowest Landau subband due to magnetic breakdown.

to magnetic breakdown beyond the CDW gap.
(2) Electric Field Effect on Lebed Resonance
We have found a novel electric field effect on angulardependent magnetotransport in Q1D layered conductors
with a pair of sheetlike Fermi surfaces. The Lebed resonance has known as the resonant increase of interlayer conduction at magnetic field orientations where electron orbits
are periodic on Fermi sheets. Under additional interlayerelectric fields, electron orbits on two Fermi sheets become
periodic at different magnetic field orientations. This causes
doubly splitting of the Lebed resonance of interlayer conduction. This effect provides us a new experimental tool toestimate the Fermi velocity of Q1D conductors.
We have successfully demonstrated this effect in an

Fig. 2. Angular dependence of interlayer conduction in α-(BEDTTTF)2KHg(SCN)4 under several interlayer electric fields at T=1.7 K.
The magnetic field with fixed strength B=13 T is rotated in the plane
normal to conducting layers with azimuth ϕ=54deg from the 1D axis.
The Lebed resonance peaks at low electric fields indicated by red
arrows split to double peaks as the electric field is increased.

New Heavy Fermion Superconductor
Ce2CoIn8
Y. Uwatoko
A new family of Ce based heavy fermion compoundswith general formula CenTIn3n+2 (T = Co, Rh, Ir; n = ∞, 1,2) show very interesting physical phenomena such as a new
type of unconventional superconductivity at ambient pressure and pressure induced superconductivity emerging upon
the suppression of an antiferromagnetic order [1,2]. We
have found favorable conditions for the growth of single
crystal Ce2CoIn8. Measurements of specific heat and electrical resistivity were carried out on single crystals of
Ce2CoIn8. The resistivity measurement reveals that
Ce2CoIn8 is a heavy fermion superconductor with Tc ~ 0.4
K [3].
The specific heat divided by temperature C/T and theelectrical resistivity ρ for Ce2CoIn8 are shown in Fig. 1. At
Tc we observed a distinct jump in C/T, providing clear evidence of bulk superconductivity. However, in the C/T data
we could not find any evidence for the phase transitions at
Ta1 and Ta2 where ρ(T) shows a substantial reduction. C/T of
Ce2CoIn8 exhibits a Schottky-like peak at around 0.6 K and
2 K. The estimated value of ∆C/γTc ~ 0.22 suggests very
weak coupling. Here the γ value is determined to be ~ 460
mJ/K2mol-Ce from the value of C/T just above Tc and ∆C is
the jump in the specific heat at Tc. This ratio, of ∆C/γTc, isapproximately equal to that of UGe2 in the superconducting
regime at 1.15 GPa [4].
The effect of a magnetic field on the resistivity of
Ce2CoIn8 was examined under various fixed fields up to 9 T
applied parallel to the a-axis. The magnetic fields in excess
of 1 T completely suppress the resistivity drop due to super-Fig. 1. Temperature dependence of the specific heat divided by temperature (red dots) and the electrical resistivity (blue dots) in thesame sample of Ce2CoIn8. The arrows show the characteristic temperatures Ta1, Ta2 and Tc determined by a peak position of the resistivity derivative dρ/dT.

ISSP Activity Report 2002

19

Fig. 2. Critical-field vs. temperature phase diagram for Ce2CoIn8
along (a) the a-axis and (b) the c-axis. Solid lines and gray mask are
guides to the eyes.

conducting transition. Ta1 and Ta2 shifted to low temperature
with increasing field. There was no structure in the resistivity at low temperature above 9 T. We have obtained the characteristic fields Hc2, Ha1 and Ha2 from the ρ(T) measurements in magnetic fields applied along the a- and the c-axes.
The results are plotted in Fig. 2 as a function of the temperature.
Since no anomaly is observed in C/T at Ta1 and Ta2, theresistivity anomalies at these temperatures would presumably be due to the nonuniformity of the sample; e.g. stacking
fault ascribable to the unique crystal structure. However, thebehavior at Ta1 and Ta2 might also originate from the quasitwo-dimensional electronic structure peculiar to the
CenTIn3n+2 families. The problem would be settled by growing a high-quality crystals and measuring low temperature
specific heat under magnetic fields.
We conclude that the transition around 0.47 K comesfrom intrinsic bulk superconductivity in Ce2CoIn8.
References[1] C. Petrovic, P. G.Pagliuso, M. F. Hundley, R.Movshovich, J. L.
Sarrao, J. D. Thompson, Z. Fisk and P.Monthoux, J. Phys. Condens.
Mater 13, L337 (2001).
[2] J. D. Thompson, R. Movshovich, Z. Fisk, F. Bouquet, N. J. Curro,
R. A. Fisher, P. C. Hammel, H. Hegger, M. F. Hundley, M. Jaime, P.
G. Pagliuso, C. Petrovic, N. E. Phillips and J. L. Sarrao, J. Magn.
Magn. Mater 226-230, 5 (2001).
[3] G. Chen, S. Ohara, M. Hedo, Y. Uwatoko, K. Saito, M. Sorai and
I. Sakamoto, J. Phys. Soc. Jpn. 71, 2836 (2002).
[4] N. Tateiwa, T. C. Kobayashi, K. Amaya, Y. Haga, R. Settai and Y.
Onuki, Physica B 312-313, 109 (2002).

most ideal two-dimensional (2D) quantum spin systems with
nuclear spin S= 1/2 on a triangular lattice. Because of thehard-core potential between 3He atoms, the higher ordermultiple spin exchange (MSE) processes as well as two-particle exchange play important roles. The exchange of evennumber of particles is antiferromagnetic (AFM), while thatof an odd ones is ferromagnetic (FM). The competition
between them, in addition to the geometrical frustration
inherent to a triangular lattice structure, makes the present
system strongly frustrated. Recent theoretical studies in theexact diagonalization method for a finite system show thatthe system favors the quantum spin liquid (QSL) with an
excitation gap.
Experimentally, the magnetizations for the 4/7 AFM
solid 3He phase adsorbed on graphite preplated by one layerof 4He and bilayer of HD have been measured above
100 µK, without any clear evidence of a spin gap. Hence theground state of the 4/7 phase is still an open question. Herethe first direct demagnetization of the 2D 3He down to10 µK region is reported.
We have prepared the three different samples: (I) the
paramagnetic solid 3He formed on graphite as the first layer,(II) the antiferromagnetic solid 3He adsorbed on one layer of
4He preplated graphite, (III) the antiferromagnetic solid 3He
adsorbed on two layers of HD preplated graphite. They weredemagnetized slowly by themselves in 17 hours from theinitial condition of Bi=0.55 T and Ti=150 µK to the final
field Bf= 5mT or 2.5 mT, where NMR measurements havebeen made to clarify their magnetization. The result fromthe sample (I) is essential to know the demagnetization
process, because the sample temperature can be determinedeasily from the magnetization which obeys a Brillouin function. According to the results, the sample is rather strongly
connected with the Cu nucei on the second stage and theheat bath model was proposed to explain the process. Basedon it, the demagnetized temperatures of the AFM samples(II) and (III) were determined.
The magnetizations for the antiferromagnetic solid 3He
adsorbed on both one layer of 4He and two layers of HD
preplated graphite are shown in Figure 1. The data in thehigh temperature region are also shown there. They increase
gradually down to about 10 µK. However no evidence corresponding to a spin gap was observed even for at much
lower temperatures than the exchange interaction by a factor
3

4

3

4

He/ He/Gr(Demagnetization)
He/ He/Gr(Thermal equilibrium)

3

He/HD/HD/Gr(Demagnetization)

3

He/HD/HD/Gr(Thermal equilibrium)

Authors
M. Hedo, N. Kurita, Y. Uwatoko, G. Chena, S. Oharaa, I. Sakamotoa
aNagoya Institute of Technology.

Gapless Spin Liquid in the Two
Dimensional Antiferromagnetic Solid 3He

Magnetization (arb. units)31
θ = -0.9 mK

0.1
θ = -9.0 mKJχ = θ / 3

H. Ishimoto
Highly frustrated low-dimensional quantum antiferromagnets have attracted much attention because a magnetically disordered ground state may be realized. A low-density
solid 3He film adsorbed on graphite surface is one of the

20

ISSP Activity Report 2002

He/HD/HD/Gr(Ikegami et al.)

0.01
-5
10

0.00010.001
0.01
Temperature (K)0.1

Fig.1. Temperature dependence of magnetization for 2D antiferromagnetic solid 3He on mono layer of 4He and bilayer of HD preplated graphite .

of almost 300, indicating a strong possibility of a quantum
spin liquid ground state with zero spin gap in the 2D triangular antiferromagnetic system with the multiple spin
exchange. These results are consistent with the heat capacity
data observed in the second layer, but in contradiction withthe theoretical prediction in the exact diagonalization for a
finite size. The system size in these calculations might be
too small because the higher order multiple spin exchange
processes play important roles. Further theoretical calculations for a larger size of the system are eagerly desired.
Authors
Ryuichi Masutomi , Yoshitomo Karaki and Hidehiko Ishimoto

Calculation of the Phase Field of a Vortex
Pair on the Surface of a Multiply
Connected Substrate
M. Kubota
The superfluid properties of thin 4He films adsorbed on
three-dimensionally (3D) connected porous media havebeen studied for many years. Helium films adsorbed onlarge-scale flat substrates are known to exhibit the classic
2D Kosterlitz-Thouless (KT) phase transition. On the otherside, thin 4He films adsorbed on 3D connected porous media
present a number of interesting physical phenomena: dilute
Bose gas behavior for low-density 4He adsorbed in Vycor
glass; 3D critical behavior observed for a wide range of systems; and 2D and 3D crossover behavior as a function ofsystem structure and pore size. It is clear that one is notdealing with ideal 2D system; yet there remains the interesting question as to what extent vestigial aspects of 2D KT
behavior may still be present in the superfluid behavior of
the system. Although there has been a number of theoretical
trials to describe this system, none of them treated the realcomplexity of the 3D connected system, instead they avoided this through simplifying assumptions.
We extend the work of the previous authors with a more
realistic calculation of the energies for quantized vortex
pairs taking into account the details of the 3D porous structure (Fig. 1). The quantum phase φ is calculated with regard
to the position on the surface of lattice through a numericalFig. 1. The porous substrate and the development of a rod. Every rod
has the same dimensions. The parameters are following: l is the lattice constant (distance between nodes), a is the thickness of a rod,
ã=l-a the length of a rod, and the ξ and η are the separation aroundand along the rod axis, respectively. With these definitions, we define
an aspect ratio γ=l/a and the porosity P=1-(3γ -2)/γ 3. The open volume of the pores forms a complementary cubic structure with thesame lattice parameter and the complementary aspect ratio γ∗=l/ã =γ /( γ -1). The entire system has a translational symmetry in each three
directions.

Fig. 2. η vortex ring (upper) and ξ vortex ring (lower)

(a)

(b)

Fig. 3. The vortex pair energy functions. (a) The energy function
when a vortex pair expands along the η axis. The broken line is theanalytic cylinder function. (b) The energy function when a vortex
pair expands along the ξ axis. The broken line is a normal KT-type
logarithmic function. The squares in the panels (a) and (b) are the ηand ξ ring energy, respectively (see the main text).

solution of Laplace’s equation while maintaining the continuity of each rod. We place a vortex pair on a single rod anddisplace them. The position of a vortex “core” is defined on
the smallest sites of grids by a 2π phase change. The flow
field and the total energy are obtained once the phase field is
calculated. Some solutions of Laplace’s equation are shownon the colored graphs in Fig. 2. Flow field is defined of thespecial derivative of phase. There is 3D coherence of phaseand the flow field is the vortex-like configuration around apore and a rod.
We plot the energy functions with respect to the vortex
configurations in Fig. 3. The vortex pair energy dependence
on the displacement is similar to the classical vortex pairenergy dependence in the smallest displacement. Once thevortex pair are moved further, the energy exhibits the peak.
The whole process shows a thermal activation process. We
calculated the porosity dependence of the energy function
and concluded some calculation based on the simple mean
field for excitations with a thermal activation process for acertain porosity case.
Principal Publication and Authors:
T. Obata and M. Kubota, Phys. Rev.B 66, 140506 (2002).

ISSP Activity Report 200221

Micro Pressure-Cell for Measurementsof Specific Heat and Magnetization
I. Umehara, M. Hedo and Y. Uwatoko
Strongly correlated electron systems under the high pressure show a variety of physical properties. In order to understand the ground state properties of the strongly correlated
electron systems under high pressures, many different piston
cylinder type high pressure cells have been designed. (Seefor example Ref. [1]) We have developed a micro pressurecell for the measurements of specific heat and magnetization, which can be used in the commercial SQUID magnetometer. Here we report about the design of our micro pressure cell.
Figure 1 shows the schematic drawing of the micro cell.
The cylinder and upper and lower nuts of the micro pressure
cell are constructed from hardened CuBe alloy (C1720B-H).
The dimensions of the cylinder are 8.8, 2.7 and 21 mm forthe outer diameter, inner diameter and length, respectively.
ZrO2 was used for the piston and piston backups. The highpressure sample space was sealed by Teflon cell techniquewith Cu seal rings. The mixture of Fluorinert FC70:FC77=1:1 was used as the pressure transmitting media. Total
weight of this cell is about 9 g. To the best of our knowledge, our cell is one of the smallest among this type. Usingsome attachments as drawn in the figure, it can be used forthe specific heat and magnetization measurements. The
magnetization studies were performed with commercial
SQUID magnetometer and the specific heat measurements
were carried out using the conventional adiabatic method. In
preparation for the specific heat measurements, we havemeasured the heat capacity of the micro pressure-cell withthe mixture of Fluorinert under several pressures. We havealso measured the magnetization of our cell as functions oftemperatures and magnetic fields.
As an example, we show in Fig. 2 the specific heat (top)and the magnetization (bottom) as a function of temperature
below 9 K for the single crystalline CeAg under variouspressures, 0, 0.30, 0.83, and 1.37 GPa. In this experiment,
the weight of the sample is 80 mg, which is less than 1 % of
the weight of the pressure cell. However, the heat capacityof the sample at around 8 K is about 20 % of the total heat
capacity. For the magnetization measurements, a magnetic
field of 100 Oe was applied along the 〈100〉 direction in a
Cubic notation at room temperature. The Curie temperature
TC increases with pressure below around 0.6 GPa and thendecreases with increasing pressure up to 1.37 GPa. The TC

Fig. 2. The specific heat and magnetization for CeAg as a function oftemperature at various pressures.

values determined from the specific heat and the magnetization measurements are in excellent agreement with eachother and with previous resistivity measurements [3,4]. We
have succeeded in applying pressures up to 2 GPa (low temperature value) in other experiments with our cell. Detailed
analysis of our measurements for the specific heat and themagnetization in CeAg will appear elsewhere.
References[1] Y. Uwatoko et al., Rev. High Pressure Sci. Technol.7, 1508
(1998).
[2] P. Morin, J. Magn. Magn. Mater. 71, 151 (1988).
[3] M. Kurisu, J. Phys. Soc. Jpn. 56, 4064 (1987).
[4] A. Elling and J.S. Scilling, Phy. Rev. Lett. 46, 364 (1981).
Authors
Izuru Umeharaa, Masato Hedo, Yoshiya Uwatoko
aDivision of Physics, Faculty of Engineering, Yokohama National
University.

World Record Fast Rotating ULTCryostats
T. Mizusaki and M. Kubota
Superfluidity is a macroscopic quantum phenomena
where the phase of the condensate or bulk superfluid is amacroscopic physical quantity and can be controlled byexternal flow or rotational motion of the vessel. Since thefirst observation of the meniscus of the rotating superfluid
4He by Osborn, experiments under rotation brought us continued surprises [1]. This first observation led to the idea of
Onsager-Feynmann vortex lines. Henry Hall was the first to
observe persistent superflow after rotation, and W.F. Vinen
found that the flow around a wire is quantized with a circulation given by h/m, where h is Plank’s constant and m is the
4He atom mass. Since then, there have been a number ofexperiments performed under rotation for 4He. Superfluid
made of Fermi particle, 3He has been studied under rotation
since 80’s [2]. Kubota and his group have been studying
till present time a new class of 3D superfluid; submonolayer

Fig. 1. Schematic drawing of micro pressure cell.

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ISSP Activity Report 2002

[3] M.K. Zalalutdinov et al., Czechoslovak J. Phys. 46, S1-39 (1996);
M. Fukuda et al., J. Low Temp. Phys. 113, 417 (1998); M. Fukuda etal., Physica B 250-252, 117 (2000).
[4] R. Ishiguro et al., Physica B 329-333, 66 (2003).
Principal Publication and Authors
Minoru Kubota, Toshiaki Obata, Ryousuke Ishiguro, Minoru
Yamashita,Takeshi Igarashi, Emi Hayat, Osamu Ishikawa, Yutaka
Sasaki, Nikolay Mikhin; Muneyuki Fukuda, Vladimir Kovacik , Takao
Mizusaki, Physica B 329-333, Part 1, 577 (2003).

Trajectory of the Wave Packets Oscillating
on an Adiabatic Potential Surface
T. SuemotoFig. 1. ISSP ULT Rotation Cryostat with rotating scheme Tok-01.
Two high precision air bearing units are located at the ends of the
shaft. Electronics are located on a table above and driven by precision
motor through timing belt. The connection between this table and theshaft which connects rigidly the cryostat is made by a soft torsion
rod, which functions also as a rotational vibration isolator. Earth fieldcancellation coils are seen on a thick paper cylinder around the
Dewar.
4He film systems condensed on 3D connected surface ofporous substrates. And some experiments have been performed under rotation [3]. Now we are adding ourselves to
the groups of 3He study under rotation, as well, after ones atCornell, Helsinki, Berkeley, and Manchester: Ishiguro et al.
have reported vortex nucleation and textures in small cylinders [4], which indicates an interesting features of thisanisotropic superfluid made of Cooper pairs with large internal freedom.
Rotation for the study of superfluids is compared sometimes with magnetic field for the study of superconductors.
Rotational speed, for example, to destroy superfluidity canbe compared with critical magnetic field Hc2, but as to rotation, it has been far beyond reach of the present known technology to overcome Hc2 correspondance for T=0 K for bulk
superfluids. We are, however, making some progress andunder some conditions of combination of the new highspeed rotating cryostat (Fig.2) and the artificial 3D superfluid consisting of thin He film and porous media we may
come close to cover interesting phase fields of a superfluid.

Real time tracking of the movement of atoms is a fundamental desire in investigating the ultrafast lattice rearrangement phenomena in solids. In this study, we have shown thatthe real time tracking of precise spatial positions of atoms ispossible by using a luminescence method.
The self-trapped excitons (STE) are generated by 20 fs
light pulses at 1.6 eV in quasi-one-dimensional halogenbridged platinum complex [Pt(en)2] [Pt(en)2Br2](ClO4)2(en=ethylendiamine). Previously, we observed intensity
modulation of the luminescence with a period of about 290
fs [1], corresponding to the oscillations of the wave-packets
on the adiabatic potential curve of the lowest STE state (seethe black dots in Fig. 1).
We have greatly improved the experimental setup for the
up-conversion spectroscopy and performed a measurementwith a time resolution of about 40 fs, which we believe isthe best value ever achieved. At a photon energy of 0.8 eV,
which corresponds to the luminescence from the bottom of
the adiabatic potential curve, we observed the overtone
modulation with a period of about 145 fs as shown in Fig. 1
by green arrows. This has not been seen in the measurementwith 90 fs time resolution, while it was theoretically expected to be observed if the time resolution is improved [1].
Modulation frequency was doubled, because the wave-packet passes thorough the bottom of the adibatic potential curve

References[1] see for example, E.L. Andronikashvili and Yu.G. Mamaladzs, Rev.
Mod. Phys. 39, 567 (1966).
[2] O. Lounasmaa and E. Thuneberg, Proc. Nat. Acad. Sci. USA, 96(1999).

Fig. 2. ISSP High Speed Rotating Cryostat. Experiments under up to
6 rotation/sec are tested at dilution refrigerator temperatures.

Fig. 1. Time evolution of the luminescence intensity at various photon energies. At 1.1 eV, the oscillation period corresponds to that of
the wave packet motion. At 0.8 eV overtone oscillation is observed.

ISSP Activity Report 2002

23

Fig. 2. The positions of the peaks and shoulders in the time evolution
curves in Fig. 1 are shown by blue circles on the coordinates of timeand photon energy. The scale on the right side shows the approximate
displacement of the Br from the equilibrium position. The red curveshows the motion of the peak of the wave packet assuming a damped
harmonic oscillator.

twice within one period in alternative directions. We can see
also the anti-phasing behavior between the oscillations ofluminescence at the photon energies of 0.7 eV and 0.9 eV(shown by blue arrows). The luminescence at these twoenergies probes the wave-packet at the points on the opposite sides of the potential minimum. This clearly shows thatthe wave-packet is moving back and forth around the potential minimum. The peaks or the shoulders in each curve corresponds to the moment when the peak of the wave packet
passes on the point corresponding to the observed photon
energy. Thus we can make a trajectory of the wave packet
moving in the adiabatic potential curve by plotting the peak
points as shown in Fig. 2. If we assume the same curvature
for the ground state and the excited state, the luminescenceenergy can be directly translated into the position of theoscillating ion (Br in this case). Thus we succeeded in tracking the spatial positions of atoms [2].
References[1] S. Tomimoto et. al., Phys. Rev. B 60, 7961 (1999).
[2] T. Matsuoka, S. Saito, J. Takeda, S. Kurita, and T. Suemoto,Nonlinear Optics 29 (10-12), 587-593 (2002).
Authors
Taira Matsuoka, Tohru Suemoto, Jun Takedaa and Susumu Kuritaa
a Yokohama National University

Three-Dimensional Imaging of Ge Atoms
on Si(001) by Using a Holographic
Method in X-ray Diffraction

Fig. 1. Principle of the present method to reconstruct images of surface atoms using an interference effect.

surface atoms.
Figure 1 illustrates the principle of the present method.
In this case the diffraction amplitude from a sample is divided in two; one is a contribution from a surface layer to be
determined, and the other is a contribution from a substrate
crystal given by the sum of amplitudes from a semi-infinite
layers. In most cases we know the structure of the substrate
crystal, and can calculate the phase and modulus of theamplitude based on the X-ray diffraction theory. Thus we
can regard the wave scattered from the substrate crystal as areference wave. In this situation, we can reconstructunknown surface atoms by a Fourier transform after a proper normalization of the experimental data, instead of illuminating a reference wave as in optical holography.
Figure 2 shows images of Ge atoms on Si(001) substrate
crystal reconstructed from observed intensities. Here weused a sample in which a monolayer of Ge atoms are epitaxially grown on a clean Si(001) surface at 775 K. A few
nanometers of an amorphous Si layer was further depositedon the sample at 295 K to treat in the atmosphere. To reconstruct the images, we have obtained intensities about at 250
points, most of which are measured at points on the crystaltruncation rod close to Bragg points. The result clearlyshows that Ge atoms locate positions that Si atoms would
otherwise occupy. The position of Ge atoms is slightly
relaxed toward the surface, reflecting the difference in bondlengths between Ge and Si in crystals.
In the structure analysis of surfaces and interfaces, constructing an initial model is a most elaborating task. Thus

T. Takahashi
In X-ray diffraction what we can observe is X-ray intensities, and thus the phase of X-ray waves is lost in measurements. If, however, we can get the information of phases ofdiffracted X-rays, the atomic structure of materials is reconstructed from intensity data through Fourier transform.
In the present work, we show a holographic method toreconstruct images of surface and interface atoms with an
atomic resolution. Here we treat two-dimensionally ordered
atoms on a substrate crystal. Hitherto similar attempts havebeen made as to two-dimensional structures projected on thesurface. In this work, in contrast, we show the first experimental result of three-dimensional reconstructed images of

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ISSP Activity Report 2002

Fig. 2. Contour maps of Ge atoms on Si(001) substrate crystal atplanes parallel (left) and perpendicular (right) to the surface.
Rectangular regions indicate the unit cell defined so that only a single
atom is included in a plane parallel to the surface.

once the initial structure is directly obtained from experimental data, one could easily refine the structure, with for
instance, the use of conventional fitting procedures.
Combination of the present method with direct methods will
be more promising.
Principal Publication and Authors
K. Sumitani, T. Takahashi, S. Nakatani, A. Nojima, O. Sakata, Y.
Yoda, S. Koh, T. Irisawa, and Y. Shiraki; Jpn. J. Appl. Phys. 42,
L189-L191 (2003).

Discovery of Missing Link: III. The A-site
Ordered Manganese Perovskite RBaMn2O6
Y. UedaThe manganese perovskite R3+1-xA2+xMnO3 has drawn
much attention in its rich physical properties such as ferromagnetic metal (FM) generated by the double-exchangeinteraction, charge/orbital order (CO), colossal mangetoresistence and electronic phase segregation. It is now widely
accepted that these interesting phenomena are caused by thestrong correlation/competition of multi degrees of freedom,that is, spin, charge, orbital and lattice, and could be significantly influenced by the A-site randomness. Although thesearch for the compound without the A-site randomness has
been motivated by its intriguing properties, it has been missing in manganese perovskites.
Recently, the A-site ordered manganese perovskite
RBaMn2O6 (R = Y and rare earth elements) has successfully
synthesized and investigated in the structure and electromagnetic properties. The most significant structural featureof RBaMn2O6 is that the MnO2 square sublattice is sand-Fig. 1. Phase diagram for RBaMn2O6 expressed as a function of theratio of ionic radius, rR3+/rBa2+. The A-site ordered manganites
RBaMn2O6 can be classified into three groups from the structural andelectromagnetic properties. The first group (R = Y, Tb, Dy and Ho)shows successive three phase transitions, structural transition,
charge/orbital order (COI) and antiferromagnetic order (AFI), on
cooling. The second group (R = Sm, Eu and Gd) exhibits a COI transition, followed by an AFI. The third group (R = La, Pr and Nd) has aferromagnetic metal (FM) transition, followed by an A-type AF(AFM) in PrBaMn2O6 and NdBaMn2O6. The electronic phase segregation is observed in LaBaMn2O6.

wiched by two types of rock-salt layers, RO and BaO, withmuch different sizes and consequently the MnO6 octahedron
itself is heavily distorted. The electromagnetic properties are
summarized in Fig. 1 as an electronic phase diagram.
Compared with the A-site disordered R0.5A0.5MnO3 (A = Sr,
Ca), RBaMn2O6 displays remarkable features: (1) the COtransition at TCO far above 300 K, (2) a new stacking variation of the CE-type CO with a 4-fold periodicity along the caxis (4CE-CO), (3) the presence of structural transition possibly accompanied by dx2-y2 orbital order at Tt above TCO,
and (4) the electronic phase segregation. The relatively highTCO and the absence of critical region can be understood as
the effect of the absence of A-site randomness. The 4CE-CO
could originate from the layer-type order of A-site cations.
The structural transition at Tt is the most remarkable featurein RBaMn2O6. The large mismatch between RO and BaO
lattices introduces a strong frustration to the MnO2 sheet andmust give a new perturbation to the competition of multidegrees of freedom among charge, orbital, spin and lattice.
The last feature, that is, the coexistence of FM and CO phases as the ground state in LaBaMn2O6 far from CO state,might be explained as the effect of the elongation of Mn-OMn distance (reduction of bandwidth). Such effect has neverbeen seen in ordinary R3+1-xA2+xMnO3.
Principal Publication and Authors
T. Nakajima, H. Kageyama, H. Yoshizawa and Y. Ueda, J. Phys. Soc.
Jpn. 71, 2843 (2002).

Photocarrier Injection to a Transition
Metal Oxide
Z. HiroiAn efficient method to dope a transition metal oxide(TMO) with hole carriers is presented, which is photocarrier
injection (PCI) in an oxide heterostructure. It is shown thatan insulating vanadium dioxide (VO2) film is rendered
metallic under light irradiation by PCI from an n-type titanium dioxide (TiO2) substrate doped with Nb. Consequently,
large photoconductivity, which is exceptional for TMOs, isfound in the VO2/TiO2:Nb heterostructure. We propose anelectronic band structure where photo-induced holes createdin TiO2:Nb can be transferred into the filled V 3d band via
the low-lying O 2p band of VO2.
We have fabricated a 10-nm-thick VO2 film epitaxiallygrown on the (001) surface of a Nb-doped TiO2 single crystal substrate with nominal Nb concentration of 0.05 wt%.
The deposition was carried out in a vacuum chamber using apulsed laser deposition technique with a KrF excimer laser(λ=248 nm) and a V2O3 pellet as a target.
Figure 1 shows the temperature dependence of in-plane
resistance measured in the dark and under light irradiation.
In the dark, a large jump in resistance is observed at T =297 K on heating and T = 287 K on cooling due to the MI
transition: The TMI is lower by 40 K than that of a single
crystal. The thermal hysteresis is due to the first-ordernature of the phase transition. When the sample is irradiatedby an UV light (λ=300~400 nm), the resistance shows asmall jump at almost the same TMI as in the dark.
Surprisingly observed below TMI is a dramatic reduction inresistance by more than two orders of magnitude. In addition, the resistance decreases gradually with decreasing temperature, indicating a metallic behavior, in contrast to an

ISSP Activity Report 2002

25

Fig. 1. Temperature dependence of in-plane resistance measured oncooling and heating. The sample is a 10-nm-thick VO2 film grownon an n-type TiO2:Nb (nominally 0.05 wt%) substrate with 0.5 mm
thickness. An ultraviolet (UV) light (wavelength: 300-400 nm; irradiance: 140 mWcm-2) or a visible (VIS) light (wavelength: 400-700
nm; irradiance: 350 mWcm-2) from a xenon lamp was guided via an
optical fiber to irradiate the sample during the measurements.

insulating temperature dependence seen in the dark. We also
detected a small reduction of about 4 % in resistance above
TMI. On the other hand, a similar but smaller reduction isobserved for visible-light irradiation. These photoconductivity occurs quite reversible upon light switching.
We propose a simple picture for the electronic structureof the VO2/TiO2:Nb system to explain the experimental
facts qualitatively (Fig. 2). It is well established that TiO2
possesses a valence band of nearly O 2p orbital character
and a conduction band of Ti 3d character. Nb5+ donor levels
exist just below the conduction band. On the other hand,VO2 has a nearly half-filled d// band in the metallic phasewhich splits into two, d// and d//* bands, in the insulatingphase. To combine the two materials into a junction, assuming Fermi level matching, results in almost uniquely theenergy diagrams shown in Fig. 2 for two situations; one athigh temperature above TMI in the dark and the other at low
temperature below TMI under UV light irradiation. It is to be
noted that the band structure of the junction would be completely different without donors in TiO2, because the Fermi

Fig. 2. Schematic drawings of photocarrier injection (PCI) from a
TiO2:Nb substrate to a VO2 film under light irradiation and of theelectronic structure for the VO2/TiO2:Nb heterostructure.

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ISSP Activity Report 2002

level would lie in the middle of the gap of TiO2 in that case.
Light absorption should occur mostly in the thick (0.5mm) TiO2:Nb crystal, creating holes and electrons in thevalence and conduction bands, respectively. Then, it is naturally expected from the energy diagram for the low-temperature case that only the holes can migrate into the O 2p bandof VO2, because the top of the O 2p band of VO2 is locatedslightly above that of the valence band of TiO2:Nb, while
the bottom of the empty d//* band of VO2 is much above thatof the conduction band of TiO2:Nb. A possible upwardbending of the TiO2 bands near the interface may help thiselectron-hole separation. With increasing hole density inVO2, the VO2 bands should shift downward by an induced
open-circuit voltage VOC which was in fact detected in ourtransverse voltage measurement across the interface. The
transferred holes in the O 2p band of VO2 must find theircounter electrons located near the Fermi level to be combined, resulting in conducting hole carriers in the d// band at
low temperature below TMI. As a result, the VO2 film
becomes a metal, exhibiting a large photoconductivity or an
insulator-to-metal transition under light irradiation.
The present PCI method would provide us with an efficient and convenient way for tunable hole doping in many
TMOs. From the physical point of view, PCI would be very
interesting to test the effect of ‘clean’ hole doping. On theother hand, from the application point of view, the present
heterostructure could be used as a new type of optoelectronics device using an UV light.
Principal Publication and Authors
Y. Muraoka, T. Yamauchi, Y. Ueda and Z. Hiroi; J. Phys.: Condens.
Matter 14, L757 (2002).

Novel In-gap State in Zn-doped
La1.85Sr0.15CuO4
K. Hirota
It is a received wisdom that the antiferromagnetism (AF)on a hole-doped CuO2 plane in a lamellar copper oxide isrelevant to the high Tc superconductivity. Our aim is tounderstand the spin state of doped CuO2 plane, particularlyin the superconducting state. However, there exists no AF
long-range order (LRO), and the low-energy spin excitations, which might play an important role in the high Tc
superconductivity, disappear due to a so-called spin gap at
low temperatures. The best way to observe the low-energy
spin response is to suppress the superconductivity by giving
some perturbations to the spin system without changing theother properties. Zn doping is extremely suitable for the purpose: A small amount of Zn2+ ions substituting for Cu2+ions strongly suppress the superconductivity without doping
holes nor giving much structural disturbance. NMR and neutron studies have revealed that doped Zn impurities induce
staggered magnetic moments on the surrounding Cu sites,
indicating that Zn-doping enhances AF correlations on a
CuO2 plane [1,2]. If we can tune Tc by precisely controlling
the Zn concentration in single crystals large enough forinelastic neutron scattering, it would provide a chance tostudy in detail the microscopic nature of spin correlations
and their contribution to the high Tc paring mechanism.
We have carried out a comprehensive neutron scattering
study of the AF spin correlations in Zn-doped
La1.85Sr0.15CuO4 single crystals to elucidate how the spin-correlations originate not from the independent local magnetisms around Zn impurities but from the long-range AF
coherence among the induced moments around different Zn
ions. Thus we conclude that the in-gap state continuously
connects with an AF ground state with increasing Zn, i.e.,decreasing RZn-Zn, indicating the importance of underlying
AF ground state which is locally substituted for the superconducting state with help of small perturbations.
References[1] M.-H. Julien et al., Phys. Rev. Lett. 84, 3422 (2000).
[2] K. Hirota, Physica C 357-360, 82 (2001).
Authors
H. Kimuraa, M. Kofu and K. Hirota
aIMRAM, Tohoku University

Measurement of an Electron Beam Sizewith a Beam Profile Monitor Using
Fresnel Zone Plates
Fig. 1. Energy dependence of the q-integrated χ''(ω) for (a) y=0.004,(b) y=0.008 and (c) y=0.017 at 10 K. Open diamonds denote the datafor y=0.0. Solid lines are fits with a phenomelogical dynamical spin
susceptibility.

gap state is broken and how the static AF correlations are
induced by Zn-doping. To obtain quantitative information
about the Zn-doping dependence of spin excitation spectra,
it is essential to control the Zn-doping rate accurately.
Furthermore, large and spatially homogeneous crystals are
required. We have overcome such difficulties by combining
the traveling-solvent-floating-zone (TSFZ) method and aquantitative analysis of elements using the inductively-coupled plasma method. Our systematic studies with changing
the Zn-doping rate under unified experimental conditions
revealed a novel low-energy spin excitation which isinduced within the spin gap state by doped Zn impurities.
Figure 1 shows the energy spectra of the dynamical spin
susceptibility χ''(ω) around 10 K. All the data are correctedso as to be directly compared. It is remarkable that χ''(ω) for
all the samples have a maximum and almost identical intensity around ω = 8 meV, while χ''(ω) below 8 meV develops
with increasing doped Zn. These systematic changes cannot
be explained by either the broadening of the gap structure or
the reduction of the gap-energy because both the cases
should be associated with the variation of χ''(ω) near the gap
energy. Therefore, these energy spectra indicate that Zndoping induces an additional spin excitation in the spin gap
below ω~8 meV and that, with further doping, the novel
spin excitation is enhanced and shifts to lower energies, i.e.,
becomes more static. Indeed, the elastic magnetic signals
were observed for y=0.017 below ~20 K at the same incommensurate positions as that of the inelastic scattering, while
no signal was detected in y=0.008 down to 1.5 K. The inplane spin correlation length for y=0.017 is estimated at~80 Å.
A mean distance between Zn ions (RZn-Zn) shortens from~60 Å in y=0.004 to ~42 Å in y=0.008. We thus speculatethat the induced local magnetic moments around a doped Zn
ion start correlating with those around other Zn ions for
y=0.008, and that the correlations among the moments
around different Zn ions become coherent, which gives riseto the novel in-gap spin state at particular q positions. The
in-plane static spin correlation lengths for y=0.014 and0.017 exceed 80 Å, which is much longer than their RZn-Zn
values, 32 and 29 Å. This fact indicates that the static spin

N. Nakamura
In low-emittance accelerators such as third-generationsynchrotron light sources and linear-collider damping rings,
the beam size reaches less than ten micrometers. A beamprofile monitor with an x-ray imaging optics for synchrotron
radiation (SR) is expected to measure such a small electronbeam size with a high spatial resolution in a non-destructive
manner. A Fresnel zone plate (FZP) begins to be used as an
x-ray imaging lens and its spatial resolution, which depends
on the outermost-zone width of the FZP, is capable of being
less than 100 nm. A real-time beam profile monitor based on
FZPs was already proposed and designed for the Super-SOR
light source [1]. We have recently constructed the same typeof beam profile monitor in the KEK-ATF damping ring with
an emittance of about 1 nm to study the feasibility of themonitor.
Figure 1 shows the layout of the beam profile monitor
system constructed in the KEK-ATF damping ring. The
monitor system has a structure of a long-distance x-ray
microscope, where two FZPs (CZP: Condenser Zone Plateand MZP: Micro Zone Plate) constitute an x-ray imaging
optics. The synchrotron radiation from the electron beam at
the bending magnet is monochromatized by a Si(220) crystal monochromator and the electron beam profile is twentytimes magnified by two FZPs and detected on an x-ray CCD
camera. The spatial resolution is about 1.7 µm for the selected photon energy of 3.235 keV. The x-ray CCD camera is adirect incident type with a back-thinned illuminated CCD,
which offers high quantum efficiency, more than 90 % for
the selected x-ray energy.

Fig. 1. Layout of the beam profile monitor system.

ISSP Activity Report 2002

27

Fig. 2. Electron beam image obtained by the beam profile monitor.
The horizontal and vertical bars of 50 µm reflect the dimension at the
SR source point.

Fig. 3. (a) horizontal and (b) vertical beam profiles with their fitted
gaussian curves. The axes of abscissa use the dimension at the SR
source point.

Figure 2 shows a typical electron-beam image measured
by the beam profile monitor. The horizontal and verticalsizes were obtained by fitting the horizontal and verticalprofiles of the beam image to gaussian curves. Figure 3
shows the horizontal and vertical beam profiles and their fitted gaussian curves. The beam sizes were measured on fivedifferent days with different beam currents and tuning conditions of the ring. The obtained horizontal beam sizes wereabout 40 mm and the vertical beam sizes 7 - 10 µm. With
this monitor, we have succeeded in obtaining a clear electron-beam image and measuring the electron-beam size lessthan 10 µm. The good performance and prospects of themonitor were confirmed in the experiments.
Reference[1] N. Nakamura et al., Proc. of the 2001 Particle Accelerator
Conference, Chicago, 2001, p. 1628-1630.
Authors
N. Nakamura, K. Iida, H. Sakai, K. Shinoe, H. Takaki, M. Fujisawa,
H. Hayanoa, M. Nomuraa, Y. Kamiyaa, T. Kosekib, Y. Amemiyac, N.
Aokid, K. Nakayamad
aHigh Energy Accelerator Research Organization(KEK)
bInstitute of Physical and Chemical Research(RIKEN)
cGraduate School of Frontier Sciences, University of Tokyo
dToshiba Corporation

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ISSP Activity Report 2002